Statistical Physics
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Claudine Hermann

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Statistical Physics

Advanced Texts in Physics This program of advanced texts covers a broad spectrum of topics that are of current and emerging interest in physics. Each book provides a comprehensive and yet accessible introduction to a field at the forefront of modern research. As such, these texts are intended for senior undergraduate and graduate students at the M.S. and Ph.D. levels; however, research scientists seeking an introduction to particular areas of physics will also benefit from the titles in this collection.

Claudine Hermann

Statistical Physics Including Applications to Condensed Matter

With 63 Figures

Claudine Hermann Laboratoire de Physique de la Matie` re Condense´ e Ecole Polytechnique 91128 Palaiseau France [email protected]

Library of Congress Cataloging-in-Publication Data is available.

ISBN 0-387-22660-5

Printed on acid-free paper.

© 2005 Springer Science+Business Media, Inc. All rights reserved. This work may not be translated or copied in whole or in part without the written permission of the publisher (Springer Science+Business Media, Inc., 233 Spring Street, New York, NY 10013, USA), except for brief excerpts in connection with reviews or scholarly analysis. Use in connection with any form of information storage and retrieval, electronic adaptation, computer software, or by similar or dissimilar methodology now known or hereafter developed is forbidden. The use in this publication of trade names, trademarks, service marks, and similar terms, even if they are not identified as such, is not to be taken as an expression of opinion as to whether or not they are subject to proprietary rights. Printed in the United States of America. 9 8 7 6 5 4 3 2 1 springeronline.com

SPIN 10955284

(MPY)

Table of Contents Introduction

xi

Glossary

xv

1 Statistical Description of Large Systems. Postulates 1.1 Classical or Quantum Evolution of a Particle ; Phase Space 1.1.1 Classical Evolution . . . . . . . . . . . . . . . . . . . 1.1.2 Quantum Evolution . . . . . . . . . . . . . . . . . . 1.1.3 Uncertainty Principle and Phase Space . . . . . . . . 1.1.4 Other Degrees of Freedom . . . . . . . . . . . . . . . 1.2 Classical Probability Density ; Quantum Density Operator . 1.2.1 Statistical Approach for Macroscopic Systems . . . . 1.2.2 Classical Probability Density . . . . . . . . . . . . . 1.2.3 Density Operator in Quantum Mechanics . . . . . . 1.3 Statistical Postulates ; Equiprobability . . . . . . . . . . . . 1.3.1 Microstate, Macrostate . . . . . . . . . . . . . . . . 1.3.2 Time Average and Ensemble Average . . . . . . . . 1.3.3 Equiprobability . . . . . . . . . . . . . . . . . . . . . 1.4 General Properties of the Statistical Entropy . . . . . . . . 1.4.1 The Boltzmann Deﬁnition . . . . . . . . . . . . . . 1.4.2 The Gibbs Deﬁnition . . . . . . . . . . . . . . . . . . 1.4.3 The Shannon Deﬁnition of Information . . . . . . . Summary of Chapter 1 . . . . . . . . . . . . . . . . . . . . . . . . Appendix 1.1 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 1.2 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 1.3 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2 The 2.1 2.2 2.3

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Diﬀerent Statistical Ensembles. General Methods Energy States of an N -Particle System . . . . . . . . . . . . . . Isolated System in Equilibrium : “Microcanonical Ensemble” . Equilibrium Conditions for Two Systems in Contact . . . . . . 2.3.1 Equilibrium Condition : Equal β Parameters . . . . . . 2.3.2 Fluctuations of the Energy Around its Most Likely Value v

1 2 2 4 4 5 5 6 8 9 10 10 11 12 15 15 15 18 19 21 27 29 31 32 35 36 36 37

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2.4

Contact with a Heat Reservoir, “Canonical Ensemble” . . . . . 2.4.1 The Boltzmann Factor . . . . . . . . . . . . . . . . . . . 2.4.2 Energy, with Fixed Average Value . . . . . . . . . . . . 2.4.3 Partition Function Z . . . . . . . . . . . . . . . . . . . . 2.4.4 Entropy in the Canonical Ensemble . . . . . . . . . . . 2.4.5 Partition Function of a Set of Two Independent Systems 2.5 Grand Canonical Ensemble . . . . . . . . . . . . . . . . . . . . 2.5.1 Equilibrium Condition : Equality of Both T ’s and µ’s . 2.5.2 Heat Reservoir and Particles Reservoir . . . . . . . . . . 2.5.3 Grand Canonical Probability and Partition Function . . 2.5.4 Average Values . . . . . . . . . . . . . . . . . . . . . . . 2.5.5 Grand Canonical Entropy . . . . . . . . . . . . . . . . . 2.6 Other Statistical Ensembles . . . . . . . . . . . . . . . . . . . . Summary of Chapter 2 . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 2.1 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

40 41 42 44 45 46 48 49 50 51 52 53 53 55 57

3 Thermodynamics and Statistical Physics 3.1 Zeroth Law of Thermodynamics . . . . . . . . . . . . . . . . . . 3.2 First Law of Thermodynamics . . . . . . . . . . . . . . . . . . . 3.2.1 Work . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.2.2 Heat . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.2.3 Quasi-Static General Process . . . . . . . . . . . . . . . 3.3 Second Law of Thermodynamics . . . . . . . . . . . . . . . . . 3.4 Third Law of Thermodynamics . . . . . . . . . . . . . . . . . . 3.5 The Thermodynamical Potentials ; the Legendre Transformation 3.5.1 Isolated System . . . . . . . . . . . . . . . . . . . . . . . 3.5.2 Fixed N , Contact with a Heat Reservoir at T . . . . . . 3.5.3 Contact with a Heat and Particle Reservoir at T . . . . 3.5.4 Transformation of Legendre ; Other Potentials . . . . . . Summary of Chapter 3 . . . . . . . . . . . . . . . . . . . . . . . . . .

59 60 60 61 63 63 64 66 67 67 69 70 72 75

4 The Ideal Gas 4.1 Introduction . . . . . . . . . . . . . . . . . . . . . . . . . 4.2 Kinetic Approach . . . . . . . . . . . . . . . . . . . . . . 4.2.1 Scattering Cross Section, Mean Free Path . . . . 4.2.2 Kinetic Calculation of the Pressure . . . . . . . . 4.3 Classical or Quantum Statistics ? . . . . . . . . . . . . . 4.4 Classical Statistics Treatment of the Ideal Gas . . . . . 4.4.1 Calculation of the Canonical Partition Function . 4.4.2 Average Energy ; Equipartition Theorem . . . . . 4.4.3 Free Energy ; Physical Parameters (P, S, µ) . . . 4.4.4 Gibbs Paradox . . . . . . . . . . . . . . . . . . . 4.5 Conclusion . . . . . . . . . . . . . . . . . . . . . . . . . Summary of Chapter 4 . . . . . . . . . . . . . . . . . . . . . . Appendix 4.1 . . . . . . . . . . . . . . . . . . . . . . . . . . .

79 79 80 80 81 83 85 85 87 89 90 91 93 95

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vii

Appendix 4.2 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 103 5 Indistinguishability, the Pauli Principle 5.1 Introduction . . . . . . . . . . . . . . . . . . . . . 5.2 States of Two Indistinguishable Particles . . . . . 5.2.1 General Case . . . . . . . . . . . . . . . . 5.2.2 Independent Particles . . . . . . . . . . . 5.3 Pauli Principle ; Spin-Statistics Connection . . . 5.3.1 Pauli Principle ; Pauli Exclusion Principle 5.3.2 Theorem of Spin-Statistics Connection . . 5.4 Case of Two Particles of Spin 1/2 . . . . . . . . . 5.4.1 Triplet and Singlet Spin States . . . . . . 5.4.2 Wave Function of Two Spin 1/2 Particles 5.5 Special Case of N Independent Particles . . . . . 5.5.1 Wave Function . . . . . . . . . . . . . . . 5.5.2 Occupation Numbers . . . . . . . . . . . . 5.6 Return to the Introduction Examples . . . . . . 5.6.1 Fermions Properties . . . . . . . . . . . . 5.6.2 Bosons Properties . . . . . . . . . . . . . Summary of Chapter 5 . . . . . . . . . . . . . . . . . .

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113 114 116 116 116 119 119 120 121 121 123 124 124 125 126 126 126 129

6 General Properties of the Quantum Statistics 131 6.1 Use of the Grand Canonical Ensemble . . . . . . . . . . . . . . 132 6.1.1 2 Indistinguishable Particles at T , Canonical Ensemble . 132 6.1.2 Description in the Grand Canonical Ensemble . . . . . . 133 6.2 Factorization of the Grand PartitionFunction . . . . . . . . . . 134 6.2.1 Fermions and Bosons . . . . . . . . . . . . . . . . . . . . 134 6.2.2 Fermions . . . . . . . . . . . . . . . . . . . . . . . . . . 135 6.2.3 Bosons . . . . . . . . . . . . . . . . . . . . . . . . . . . . 135 6.2.4 Chemical Potential and Number of Particles . . . . . . . 136 6.3 Average Occupation Number ;Grand Potential . . . . . . . . . . 136 6.4 Free Particle in a Box ; Density of States . . . . . . . . . . . . . 138 6.4.1 Quantum States of a Free Particle in a Box . . . . . . . 138 6.4.2 Density of States . . . . . . . . . . . . . . . . . . . . . . 143 6.5 Fermi-Dirac Distribution ; Bose-Einstein Distribution . . . . . . 147 6.6 Average Values of Physical Parameters at T . . . . . . . . . . . 148 6.7 Common Limit of the Quantum Statistics . . . . . . . . . . . . 149 6.7.1 Chemical Potential of the Ideal Gas . . . . . . . . . . . 149 6.7.2 Grand Canonical Partition Function of the Ideal Gas . . 151 Summary of Chapter 6 . . . . . . . . . . . . . . . . . . . . . . . . . . 153 7 Free Fermions Properties 155 7.1 Properties of Fermions at Zero Temperature . . . . . . . . . . . 155 7.1.1 Fermi Distribution, Fermi Energy . . . . . . . . . . . . . 155 7.1.2 Internal Energy and Pressure at Zero Temperature . . . 158

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7.1.3 Magnetic Properties. Pauli Paramagnetism . . . . . . . Properties of Fermions at Non-Zero Temperature . . . . . . . . 7.2.1 Temperature Ranges and Chemical Potential Variation . 7.2.2 Speciﬁc Heat of Fermions . . . . . . . . . . . . . . . . . 7.2.3 Thermionic Emission . . . . . . . . . . . . . . . . . . . . Summary of Chapter 7 . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 7.1 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.2

8 Elements of Bands Theory and Crystal Conductivity 8.1 What is a Solid, a Crystal ? . . . . . . . . . . . . . . . . . 8.2 The Eigenstates for the Chosen Model . . . . . . . . . . . 8.2.1 Recall : the Double Potential Well . . . . . . . . . 8.2.2 Electron on an Inﬁnite and Periodic Chain . . . . 8.2.3 Energy Bands and Bloch Functions . . . . . . . . . 8.3 The Electron States in a Crystal . . . . . . . . . . . . . . 8.3.1 Wave Packet of Bloch Waves . . . . . . . . . . . . 8.3.2 Resistance ; Mean Free Path . . . . . . . . . . . . . 8.3.3 Finite Chain, Density of States, Eﬀective Mass . . 8.4 Statistical Physics of Solids . . . . . . . . . . . . . . . . . 8.4.1 Filling of the Levels . . . . . . . . . . . . . . . . . 8.4.2 Variation of Metal Resistance versus T . . . . . . . 8.4.3 Insulators’ Conductivity Versus T ; Semiconductors 8.5 Examples of Semiconductor Devices . . . . . . . . . . . . 8.5.1 The Photocopier : Photoconductivity Properties . 8.5.2 The Solar Cell : an Illuminated p − n Junction . . 8.5.3 CD Readers : the Semiconductor Quantum Wells . Summary of Chapter 8 . . . . . . . . . . . . . . . . . . . . . . .

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159 160 160 163 165 169 171 175 176 177 177 180 182 184 184 184 185 189 189 190 191 196 196 196 197 199

9 Bosons : Helium 4, Photons,Thermal Radiation 201 9.1 Material Particles . . . . . . . . . . . . . . . . . . . . . . . . . . 201 9.1.1 Thermodynamics of the Boson Gas . . . . . . . . . . . . 202 9.1.2 Bose-Einstein Condensation . . . . . . . . . . . . . . . . 203 9.2 Bose-Einstein Distribution of Photons . . . . . . . . . . . . . . 206 9.2.1 Description of the Thermal Radiation ; the Photons . . 206 9.2.2 Statistics of Photons, Bosons in Non-Conserved Number 207 9.2.3 Black Body Deﬁnition and Spectrum . . . . . . . . . . . 210 9.2.4 Microscopic Interpretation . . . . . . . . . . . . . . . . . 212 9.2.5 Photometric Measurements : Deﬁnitions . . . . . . . . . 214 9.2.6 Radiative Balances . . . . . . . . . . . . . . . . . . . . . 217 9.2.7 Greenhouse Eﬀect . . . . . . . . . . . . . . . . . . . . . 218 Summary of Chapter 9 . . . . . . . . . . . . . . . . . . . . . . . . . . 221 Solving Exercises and Problems of Statistical Physics

223

Units and Physical Constants

225

Table of Contents

ix

A few useful formulae

229

Exercises and Problems Ex. 2000 : Electrostatic Screening . . . . . . . . . . Ex. 2001 : Magnetic Susceptibility of a “Quasi-1D” Ex. 2002 : Entropies of the HC Molecule . . . . . Pr. 2001 : Quantum Boxes and Optoelectronics . . Pr. 2002 : Physical Foundations of Spintronics . .

. . . . . . . Conductor . . . . . . . . . . . . . . . . . . . . .

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231 233 234 236 237 245

Solution of the Exercises and Problems Ex. 2000 : Electrostatic Screening . . . . . . . . . . Ex. 2001 : Magnetic Susceptibility of a “Quasi-1D” Ex. 2002 : Entropies of the HC Molecule . . . . . Pr. 2001 : Quantum Boxes and Optoelectronics . . Pr. 2002 : Physical Foundations of Spintronics . .

. . . . . . . Conductor . . . . . . . . . . . . . . . . . . . . .

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251 253 255 258 259 266

Index

275

Introduction A glosssary at the end of this introduction deﬁnes the terms speciﬁc to Statistical Physics. In the text, these terms are marked by an asterisk in exponent. A course of Quantum Mechanics, like the one taught at Ecole Polytechnique, is devoted to the description of the state of an individual particle, or possibly of a few ones. Conversely, the topic of this book will be the study of systems ∗ containing very many particles, of the order of the Avogadro number N , for example the molecules in a gas, the components of a chemical reaction, the adsorption sites for a gas on a surface, the electrons of a solid. You certainly previously studied this type of system, using Thermodynamics which is ruled by “exact” laws, such as the ideal gas one. Its physical parameters, that can be measured in experiments, are macroscopic quantities like its pressure, volume, temperature, magnetization, etc. It is now well-known that the correct microscopic description of the state of a system, or of its evolution, requires the Quantum Mechanics approach and the solution of the Schroedinger equation, but how can this equation be solved when such a huge number of particles comes into play ? Printing on a listing the positions and velocities of the N molecules of a gas would take a time much longer that the one elapsed since the Big Bang ! A statistical description is the only issue, which is the more justiﬁed as the studied system is larger, since the relative ﬂuctuations are then very small (Ch. 1). This course is restricted to systems in thermal equilibrium ∗ : to reach such an equilibrium it is mandatory that interaction terms should be present in the hamiltonian of the total system : even if they are weak, they allow energy exchanges between the system and its environment (for example one may be concerned by the electrons of a solid, the environment being the ions of the same solid). These interactions provide the way to equilibrium. This approach takes place during a characteristic time, the so-called “relaxation time”, with a range of values which depends on the considered system (the study of oﬀequilibrium phenomena, in particular transport phenomena, is another branch

xi

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Introduction

of the ﬁeld of Statistical Physics, which is not discussed in this book). Whether the system is in equilibrium or not, the parameters accessible to an experiment are just those of Thermodynamics. The purpose of Statistical Physics is to bridge the gap between the microscopic modeling of the system and the macroscopic physical parameters that characterize it (Ch. 2). The studied system can be found in a great many states, solutions of the Quantum Mechanics problem, which diﬀer by the values of their microscopic parameters while satisfying the same speciﬁed macroscopic physical conditions, for example a ﬁxed number of particles and a given temperature. The aim is thus to propose statistical hypotheses on the likehood for a particular microscopic state or another one to be indeed realized in these conditions : the statistical description of systems in equilibrium is based on the postulate on the quantity statistical entropy : it should be maximum consistently with the constraints on the system under study. The treated problems concern all kinds of degrees of freedom : translation, rotation, magnetization, sites occupied by the particles, etc. For example, for a system made of N spins, of given total magnetization M , from combinatory arguments one will look for all the microscopic spin conﬁgurations leading to M ; then one will make the basic hypothesis that, in the absence of additional information, all the possible conﬁgurations are equally likely (Ch. 2). Obviously it will be necessary to verify the validity of the microscopic model, as this approach must be consistent with the laws and results of Thermodynamics (Ch. 3) It is speciﬁed in the Quantum Mechanics courses that the limit of Classical Mechanics is justiﬁed when the de Broglie wavelength associated with the wavefunction is much shorter than all the characteristic dimensions of the problem. When dealing with free indistinguishable mobile particles, the characteristic length to be considered is the average distance between particles, which is thus compared to the de Broglie wavelength, or to the size of a typical wave packet, at the considered temperature. According to this criterion, among the systems including a very large number of mobile particles, all described by Quantum Mechanics, two types will thus be distinguished, and this will lead to consequences on their Statistical Physics properties : – in dilute enough systems the wave packets associated with two neighboring particles do not overlap. In such systems, the possible potential energy of a particle expresses an external ﬁeld force (for example gravity) or its conﬁnement in a ﬁnite volume. This system constitutes THE “ideal gas” : this means that all diluted systems of mobile particles, subjected to the same external potential, have the same properties, except for those related to the particle mass (Ch. 4) ;

Introduction

xiii

– in the opposite case of a high density of particles, like the atoms of a liquid or the electrons of a solid, the wave packets of neighboring particles do overlap. Quantum Mechanics analyzes this latter situation through the Pauli principle : you certainly learnt a special case of it, the Pauli exclusion principle, which applies to the ﬁlling of atomic levels by electrons in Chemistry. The general expression of the Pauli principle (Ch. 5) speciﬁes the conditions on the symmetry of the wavefunction for N identical particles. There are only two possibilities and to each of them is associated a type of particle : - on the one hand, the fermions, such that only a single fermion can be in a given quantum state ; - on the other hand, the bosons, which can be in unlimited number in a determined quantum state. The consequences of the Pauli principle on the statistical treatment of noninteracting indistinguishable particles, i.e., the two types of Quantum Statistics, that of Fermi-Dirac and that of Bose-Einstein, are ﬁrst presented in very general terms in Ch. 6. In the second part of this course (Ch. 7 and the following chapters), examples of systems following Quantum Statistics are treated in detail. They are very important for the physics and technology of today. Some properties of fermions are presented using the example of electrons in metallic (Ch. 7) or, more generally, crystalline (Ch. 8) solids. Massive boson particles, in conserved number, are studied using the examples of the superﬂuid helium and the Bose-Einstein condensation of atoms. Finally, the thermal radiation, an example of a system of bosons in non-conserved number, will introduce us into very practical current problems (Ch. 9). A topic in physics can only be fully understood after suﬃcient practice. A selection of exercises and problems with their solution is presented at the end of the book. This introductory course of Statistical Physics emphasizes the microscopic interpretation of results obtained in the framework of Thermodynamics and illustrates its approach, as much as possible, through practical examples : thus the Quantum Statistics will provide an opportunity to understand what is an insulator, a metal, a semiconductor, or what is the principle of the greenhouse eﬀect that could deeply inﬂuence our life on earth (and particularly that of our descendants !). The content of this book is inﬂuenced by the previous courses of Statistical Physics taught at Ecole Polytechnique : the one by Roger Balian From microphysics to macrophysics : methods and application to statistical physics, volume I translated by D. ter Haar and J.F. Gregg, volume II translated by D.

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Introduction

ter Haar, Springer Verlag Berlin (1991), given during the 1980s ; the course by Edouard Brézin during the 1990s. I thank them here for all that they brought to me in the stimulating ﬁeld of Statistical Physics. Several discussions in the present book are inspired from the course by F. Reif Fundamentals of Statistical and Thermal Physics, Mac Graw-Hill (1965), a not so recent work, but with a clear and practical approach, that should suit students attracted by the “physical” aspect of arguments. Many other Physical Statistics text books, of introductory or advanced level, are edited by Springer. This one oﬀers the point of view of a top French scientiﬁc Grande Ecole on the subject. The present course is the result of a collective work of the Physics Department of Ecole Polytechnique. I thank the colleagues with whom I worked the past years, L. Auvray, G. Bastard, C. Bachas, B. Duplantier, A. Georges, T. Jolicoeur, M. Mézard, D. Quéré, J-C. Tolédano, and particularly my coworkers from the course of Statistical Physics “A”, F. Albenque, I. Antoniadis, U. Bockelmann, J.-M. Gérard, C. Kopper, J.-Y. Marzin, who brought their suggestions to this book and with whom it is a real pleasure to teach. Finally, I would like to thank M. Digot, M. Maguer and D. Toustou, from the Printing Oﬃce of Ecole Polytechnique, for their expert and good-humored help in the preparation of the book.

Glossary We begin by recalling some deﬁnitions in Thermodynamics that will be very useful in the following. For convenience, we will also list the main deﬁnitions of Statistical Physics, introduced in the next chapters of this book. This section is much inspired by chapter 1, The Language of Thermodynamics, from the book Thermodynamique, by J.-P. Faroux and J. Renault (Dunod, 1997). Some deﬁnitions of Thermodynamics : The system is the object under study ; it can be of microscopic or macroscopic size. It is distinguished from the rest of the Universe, called the surroundings. A system is isolated if it does not exchange anything (in particular energy, particles) with its surroundings. The parameters (or state variables) are independent quantities which deﬁne the macroscopic state of the system : their nature can be mechanical (pressure, volume), electrical (charge, potential), thermal (temperature, entropy), etc. If these parameters take the same value at any point of the system, the system is homogeneous. The external parameters are independent parameters (volume, electrical or magnetic ﬁeld) which can be controlled from the outside and imposed to the system, to the experimental accuracy. The internal parameters cannot be controlled and may ﬂuctuate ; this is the case for example of the local repartition of density under an external constraint. The internal parameters adjust themselves under the eﬀect of a modiﬁcation of the external parameters. A system is in equilibrium when all its internal variables remain constant in time and, in the case of a system which is not isolated, if it has no exchange with its surroundings : that is, there is no exchange of energy, of electric charges, of particles. In Thermodynamics it is assumed that any system, submitted to constant and uniform external conditions, evolves toward an equilibrium state that it can no longer spontaneously leave afterward. The thermal equilibrium between two systems is realized after exchanges between themselves : this is not possible if the walls which separate them are adiabatical, i.e., they do not transmit any energy. xv

xvi

Glossary

In a homogeneous system, the intensive parameters, such as its temperature, pressure, the diﬀerence in electrical potential, do not vary when the system volume increases. On the other hand, the extensive parameters such as the volume, the internal energy, the electrical charge are proportional to the volume. Any evolution of the system from one state to another one is called a process or transformation. An inﬁnitesimal process corresponds to an inﬁnitely small variation of the external parameters between the initial state and the ﬁnal state of the system. A reversible transformation takes place through a continuous set of equilibrium intermediate states, for both the system and the surroundings, i.e., all the parameters deﬁning the system state vary continuously : it is then possible to vary these parameters in the reverse direction and return to the initial state. A transformation which does not obey this deﬁnition is said to be irreversible. In a quasi-static transformation, at any time the system is in internal equilibrium and its internal parameters are continuously deﬁned. Contrarly to a reversible process, this does not imply anything about the surroundings but only means that the process is slow enough with respect to the characteristic relaxation time of the system. Some deﬁnitions of Statistical Physics : A conﬁguration deﬁned by the data of the microscopic physical parameters, given by Quantum or Classical Mechanics, is a microstate. A conﬁguration deﬁned by the data of the macroscopic physical parameters, given by Thermodynamics, is a macrostate. An ensemble average is performed at a given time on an assembly of systems of the same type, prepared in the same macroscopic conditions. In the microcanonical ensemble, each of these systems is isolated and its energy E is ﬁxed, that is, it is lying in the range between E and E + δE. It contains a ﬁxed number N of particles. In the canonical ensemble, each system is in thermal contact with a large system, a heat reservoir, which imposes its temperature T ; the energy of each system is diﬀerent but for macroscopic systems the average E is deﬁned with very good accuracy. Each system contains N particles. In the canonical ensemble the partition function is the quantity that norms the probabilities, which are Boltzmann factors. (see Ch. 2) In the grand canonical ensemble, each system is in thermal contact with a heat

Glossary

xvii

reservoir which imposes its temperature T , the energy of each system being diﬀerent. The energies of the various systems are spread around an average value, deﬁned with high accuracy in the case of a macroscopic system. Each system is also in contact with a particle reservoir, which imposes its chemical potential : the number of particles diﬀers according to the system, the average value N is deﬁned with high accuracy for a macroscopic system.

Chapter 1

Statistical Description of Large Systems. Postulates of Statistical Physics The aim of Statistical Physics is to bridge the gap between the microscopic and macroscopic worlds. Its ﬁrst step consists in stating hypotheses about the microscopic behavior of the particles of a macroscopic system, i.e., with characteristic dimensions very large with respect to atomic distances ; the objective is then the prediction of macroscopic properties, which can be measured in experiments. The system under study may be a gas, a solid, etc., i.e., of physical or chemical or biological nature, and the measurements may deal with thermal, electrical, magnetic, chemical, properties. In the present chapter we ﬁrst choose a microscopic description, either through Classical or Quantum Mechanics, of an individual particle and its degrees of freedom. The phase space is introduced, in which the time evolution of such a particle is described by a trajectory : in the quantum case, this trajectory is deﬁned with a limited resolution because of the Heisenberg uncertainty principle (§ 1.1). Such a description is then generalized to the case of the very many particles in a macroscopic system : the complexity arising from the large number of particles will be suggested from the example of molecules in a gas (§ 1.2). For so large numbers, only a statistical description can be considered and § 1.3 presents the basic postulate of Statistical Physics and the concept of statistical entropy, as introduced by Ludwig Boltzmann (1844-1906), an Austrian physicist, at the end of the 19th century.

1

2

1.1

Chapter 1. Statistical Description of Large Systems

Classical or Quantum Evolution of a Particle ; Phase Space

For a single particle we describe the time evolution ﬁrst in a classical framework, then in a quantum description. In both cases, this evolution is conveniently described in the phase space.

1.1.1

Classical Evolution px

t1 t0

O

x

Fig. 1.1 : Trajectory of a particle in the phase space. Consider a single classical particle, of momentum p0 , located at the coordinate r0 at time t0 . It is submitted to a force F0 (t). Its time evolution can be predicted through the Fundamental Principle of Dynamics, since d p = F (t) dt

(1.1)

This evolution is deterministic since the set (r, p) can be deduced at any later time t. One introduces the one-particle “phase space”, at six dimensions, of coordinates (x, y, z, px , py , pz ). The data (r, p) correspond to a given point of this space, the time evolution of the particle deﬁnes a trajectory, schematized on the Fig. 1.1 in the case of a one-dimension space motion. In the particular case of a periodic motion, this trajectory is closed since after a period the particle returns at the same position with the same momentum : for example the abscissa x and momentum px of a one-dimension harmonic oscillator of

Classical or Quantum Evolution of a Particle

3

mass m and frequency ω are linked by : 1 p2x + mω 2 x2 = E (1.2) 2m 2 In the phase space this relation is the equation of an ellipse, a closed trajectory periodically described (Fig. 1.2). Another way of determining the time px

t0 t1 O

E

x

Fig. 1.2: Trajectory in the phase space for a one-dimension harmonic oscillator of energy E. evolution of a particle is to use the Hamilton equations (see for example the course of Quantum Mechanics by J.-L. Basdevant and J. Dalibard, including a CDROM, Springer, 2002) : the motion is deduced from the position r and its corresponding momentum p. The hamiltonian function associated with the total energy of the particle of mass m is given by : p2 + V (r) (1.3) 2m where the ﬁrst term is the particle kinetic energy and V is its potential energy, from which the force in (1.1) is derived. The Hamilton equations of motion are given by : ⎧ ∂h ⎪ ⎨ r˙ = ∂ p (1.4) ⎪ ⎩ p˙ = − ∂h ∂r and are equivalent to the Fundamental Relation of Dynamics (1.1). h=

Note : Statistical Physics also applies to relativistic particles which have a diﬀerent energy expression.

4

1.1.2

Chapter 1. Statistical Description of Large Systems

Quantum Evolution

It is well known that the correct description of a particle and of its evolution requires Quantum Mechanics and that to the classical hamiltonian function h ˆ The Schroedinger equation corresponds the quantum hamiltonian operator h. provides the time evolution of the state |ψ : i

∂|ψ ˆ = h|ψ ∂t

(1.5)

The squared modulus of the spatial wave function associated with |ψ gives the probability of the location of the particle at any position and any time. Both in Classical and Quantum Mechanics, it is possible to reverse the time direction, i.e., to replace t by −t in the motion equations (1.1), (1.4) or (1.5), and thus obtain an equally acceptable solution.

1.1.3

Uncertainty Principle and Phase Space

You know that Quantum Mechanics introduces a probabilistic character, even when the particle is in a well-deﬁned state : if the particle is not in an eiˆ the measurement result is uncertain. genstate of the measured observable A, Indeed, for a single measurement the result is any of the eigenvalues aα of the considered observable. On the other hand, when this measurement is reproduced a large number of times on identical, similarly prepared systems, the ˆ average of the results is ψ|A|ψ, a well-deﬁned value. px

h

O

x

Fig. 1.3: A phase space cell, of area h for a one-dimension motion, corresponds to one quantum state.

Classical Probability Density ; Quantum Density Operator

5

In the phase space, this quantum particle has no well-deﬁned trajectory since, at a given time t, in space and momentum coordinates there is no longer an exact position, but a typical extent of the wave function inside which this particle can be found. Another way of expressing this diﬃculty is to state the Heisenberg uncertainty principle : if the particle abscissa x is known to ∆x, then the corresponding component px of its momentum cannot be determined with an accuracy better than ∆px , such that ∆x · ∆px ≥ /2, where = h/2π, h being the Planck constant. The “accuracy” on the phase space trajectory is thus limited, in other words this trajectory is “blurred” : the phase space is therefore divided into cells of area of the order of h for a one-dimension motion, h3 if the problem is in three dimensions, and the state of the considered particle cannot be deﬁned better than within such a cell (Fig. 1.3). This produces a discretization of the phase space. (In the case of a classical particle, it is not a priori obvious that such a division is necessary, and if it is the case, that the same cell area should be taken. It will be stated in § 1.2.2 that, for sake of consistency at the classical limit of the quantum treatment, this very area should be chosen.)

1.1.4

Other Degrees of Freedom

Up to now we have only considered the degrees of freedom related to the particle position, which take continuous values in Classical Mechanics and are described by the spatial part of the wave function in Quantum Mechanics. If the particle carries a magnetic moment, the wave function includes an additional component, with discrete eigenvalues (see a Quantum Mechanics course).

1.2

Classical Probability Density ; Quantum Density Operator

When considering a macroscopic system, owing to the extremely large number of parameters, the description of the particles’ evolution can only be done through a statistical approach. The tools are diﬀerent for either the classical or the quantum description of the particles’ motion.

6

1.2.1

Chapter 1. Statistical Description of Large Systems

Necessity of a Statistical Approach for Macroscopic Systems

The physical parameters referring to individual particles, the evolution of which was described in the above section, are measured at the atomic scale (distances of the order of a Bohr radius of the hydrogen atom, i.e., a tenth of a nm, energies of the order of a Rydberg or of an electron-volt, characteristic times in the range of the time between two collisions in a gas, that is about 10−10 sec for N2 in standard conditions). For an object studied in Thermodynamics or measured in usual laboratory conditions, the scales are quite diﬀerent : the dimensions are currently of the order of a mm or a cm, the energies are measured in joules, the measurement time is of the order of a second. Thus two very diﬀerent ranges are evidenced : the microscopic one, described by Classical or Quantum Mechanics laws, unchanged when reversing the direction of time ; the macroscopic range, which is often the domain of irreversible phenomena, as you learnt in previous courses. Now we will see what are the implications when shifting from the microscopic scale to the macroscopic one. For that we will ﬁrst consider the case of a few particles, then of a number of particles of the order of the Avogadro number. If the physical system under study contains only a few particles, for example an atom with several electrons, the behavior of each particle follows the Mechanics laws given above. Once the position and momentum of each particle are known at t = t0 , together with the hamiltonian determining the system evolution, in principle one can deduce r and p at any further time t (yet the calculations, which are the task of Quantum Chemistry in particular, quickly become very complex). Thus in principle the problem can be solved, but one is facing practical diﬃculties, which quickly become “impossibilities”, as soon as the system under study contains “many” particles. Indeed in Molecular Dynamics, a branch of Physics, the evolution of particle assemblies is calculated according to the Classical Mechanics laws, using extremely powerful computers. However, considering the computing capacities available today, the size of the sample is restricted to several thousands of particles at most. Consider now a macroscopic situation, like those described by Thermodynamics : take a monoatomic gas of N molecules, where N = 6.02 × 1023 is the Avogadro number, and assume that at time t0 the set of data (r0 , p0 ) is known for each particle. Without prejudging the time required for their calculation, just printing the coordinates of all these molecules at a later time t, with a printer delivering the coordinates of one molecule per second, would take 2 × 1016 years ! The information issued in one day would be for 105 molecules only. Another way to understand the enormity of such an information on N molecules is to wonder how many volumes would be needed to print it.

Classical Probability Density ; Quantum Density Operator

7

The answer is : many more than the total of books manufactured since the discovery of printing ! One thus understands the impossibility of knowing the position and momentum of each particle at any time. In fact, the measurable physical quantities (for example volume, pressure, temperature, magnetization) only rarely concern the properties of an individual molecule at a precise time : if the recent achievements of near-ﬁeld optical microscopy or atomic-force microscopy indeed allow the observation of a single molecule, and only in very speciﬁc and favorable cases, the duration of such experiments is on a “human” scale (in the range of a second). One almost always deals with physical quantities concerning an assembly of molecules (or particles) during the observation time. “It is suﬃcient for the farmer to be certain that a cloud burst and watered the ground and of no use to know the way each distinct drop fell. Take another example : everybody understands the meaning of the word “granite”, even if the shape, chemical composition of the diﬀerent crystallites, their composition ratios and their colors are not exactly known. We thus always use concepts which deal with the behavior of large scale phenomena without considering the isolated processes at the corpuscular scale.” (W. Heisenberg, “Nature in the contemporary physical science, ” from the French translation, p.42 Idées NRF, Paris (1962)). Besides, some of the observed macroscopic properties may have no microscopic equivalent : “Gibbs was the ﬁrst one to introduce a physical concept which can apply to a natural object only if our knowledge of this object is incomplete. For example if the motions and positions of all the molecules in a gas were known, speaking of the temperature of this gas would no longer retain a meaning. The temperature concept can only be used when a system is insuﬃciently determined and statistical conclusions are to be deduced from this incomplete knowledge.” (W. Heisenberg, same reference, p.45). It is impossible to collect the detailed information on all the molecules of a gas, anyway all these data would not really be useful. A more global knowledge yielding the relevant physical parameters is all that matters : a statistical approach is thus justiﬁed from a physical point of view. Statistics are directly related to probabilities, i.e., a large number of experiments have to be repeated either in time or on similar systems (§ 1.2.2). Consequently, one will have to consider an assembly (or an “ensemble”) consisting of a large number of systems, prepared in a similar way, in which the microscopic structures diﬀer and the measured values of the macroscopic parameters are not necessarily identical : the probability of occurrence of a particular “event” or a particular

8

Chapter 1. Statistical Description of Large Systems

measurement is equal to the fraction of the ensemble systems for which it occurs (see § 1.3.2).

1.2.2

Classical Probability Density

The hamiltonian is now a function of the positions and momenta of the N i , . . . p N ). particles, i.e., H(r1 , . . . ri , . . . rN , p1 , . . . p Using a description analogous to that of the one-particle phase space of § 1.1.1, the system of N classical particles now has a trajectory in a new phase space, at 6N dimensions since there are 3N space and 3N momenta coordinates. As it is impossible to exactly know the trajectory, owing to the lack of information on the coordinates of the individual particles (see § 1.2.1), one can only speak of the probability of ﬁnding the system particles at time t in the neighborhood of a given point of the phase space, of coordinates N d3ri d3 pi . (r1 , . . . ri , . . . rN , p1 , . . . pi , . . . pN ) to i=1

On this phase space, a measure dτ and a probability density D have to be deﬁned. Then the probability will be given by D(r1 , . . . , ri , . . . , p1 , . . . , pi , . . . t)dτ . From the Liouville theorem (demonstrated in Appendix 1.1) the phase space N elementary volume d3ri d3 pi is conserved for any time t during the pari=1

ticles evolution. A consequence is that for a macroscopic equilibrium, the only situation considered in the present book, the probability density is timeindependent. The elementary volume of the N -particle phase space is homogeneous to an action to the power 3N , that is, [mass × (length)2 × (time)−1 ]3N , and thus has the same dimension as h3N , the Planck constant to the same power. The inﬁnitesimal volume will always be chosen large with respect to the volume of the elementary cell of the N -particle phase space, so that it will contain a large number of such cells. (The volume of the elementary cell will be chosen as h3N , from arguments similar to the ones for a single particle.) Finally, dτ may depend on N , the particle number, through a constant CN . Here we will take dτ =

N CN 3 3 d ri d pi . The quantity dτ is thus dimensionh3N i=1

less and is related to the “number of states” inside the volume d3ri d3 pi ; the constant CN is determined in § 4.4.3 and 6.7 so that dτ should indeed give

Classical Probability Density ; Quantum Density Operator

9

the number of states in this volume : this choice has a interpretation in the quantum description, and the continuity between classical and quantum descriptions is achieved through this choice of CN . The classical probability density D(r1 , . . . ri , . . . , p1 , . . . , pi , . . . , t) is then deﬁned : it is such that D(r1 , . . . ri , . . . , p1 , . . . , pi , . . . , t) ≥ 0, and D(r1 , . . . ri , . . . , p1 , . . . pi , . . . , t)dτ = 1 (1.6) on the whole phase space. The average value at a given time t of a physical parameter A(ri , p i , t) is then calculated using : A(t) = A(ri , p i , t)D(ri , p i , t)dτ (1.7) In the same way, the standard deviation σA , representing the ﬂuctuation of the parameter A around its average value at a given t, is calculated from : 2 (σA (t))2 = A2 (t) − A(t)2 = A(ri , p i , t) − A(ri , p i , t) D(ri , p i , t)dτ (1.8)

1.2.3

Density Operator in Quantum Mechanics

By analogy with the argument of § 1.1.2, if at time t0 the total system is in the state |ψ, now with N particles, the Schroedinger equation allows one to deduce its evolution through : i

∂|ψ ˆ = H|ψ ∂t

(1.9)

ˆ is the hamiltonian of the total system with N particles. where H It was already pointed out that, when the system is in a certain state |ψ, the result of a single measurement of the observable Aˆ on this system is uncertain ˆ : the result is one of the eigenvalues (except when |ψ is an eigenvector of A) ˆ with the probability |ψ|ϕα |2 , associated with the projection of |ψ aα of A, on the eigenvector |ϕα corresponding to the eigenvalue aα (that we take here to be nondegenerate, for sake of simplicity). If the measurement is repeated ˆ many times, the average of the results is ψ|A|ψ. Now another origin of uncertainty, of a diﬀerent nature, must also be included if the state of the system at time t is already uncertain. Assume that the

10

Chapter 1. Statistical Description of Large Systems

macroscopic system may be found in diﬀerent orthonormal states |ψn , with probabilities pn (0 ≤ pn ≤ 1). In such a case the results average for repeated measurements will be : ˆ n = ˆ n = Tr (D ˆ A) ˆ ˆ = pn ψn |A|ψ pn ψn |ψn ψn |A|ψ (1.10) A n

n

The density operator, deﬁned by ˆ = D

pn |ψn ψn |

(1.11)

n

has been introduced (its properties are given in Appendix 1.2). This operator includes both our imperfect knowledge of the system’s quantum state, through the pn factors, and the ﬂuctuations related to the quantum measurement, which would already be present if the state |ψn was certain (this state appears through its projector |ψn ψn |). One veriﬁes that ˆ n = A ˆ ˆ A) ˆ = pn ψn |ψn ψn |A|ψ (1.12) Tr (D n

n

since ψn |ψn = δn n . In the current situations described in the present course, the projection states ˆ of the N -particle total |ψn will be the eigenstates of the hamiltonian H system, the eigenvalues of which are the accessible energies of the system. In ˆ will thus be diagonal on such a basis : the problems considered the operator D each energy eigenstate |ψn will simply be weighted by its probability pn . Now hypotheses are required to obtain an expression of D(ri , pi , t) in the classical description, or of the probabilities pn in the quantum one.

1.3

Statistical Postulates ; Equiprobability

We should now be convinced that a detailed microscopic treatment is impossible for a system with a very large number of particles : we will have to limit ourselves to a statistical approach. It is necessary to deﬁne the microscopic and macroscopic states to which this description will apply.

1.3.1

Microstate, Macrostate

A conﬁguration deﬁned by the data of microscopic physical parameters (for example, positions and momenta of all the particles ; quantum numbers characterizing the state of each particle ; magnetization of each paramagnetic

Statistical Hypotheses ; Equiprobability

11

atom localized in a solid) is a microstate ∗ .1 A conﬁguration deﬁned by the value of macroscopic physical parameters (total energy, pressure, temperature, total magnetization, etc.) is a macrostate ∗ . Obviously a macrostate is almost always realized by a very large number of microstates : for example, in a paramagnetic solid, permuting the values of the magnetic moments of two ﬁxed atoms creates a new microstate associated with the same macrostate. Assume that each magnetic moment can only take two values : +µ and −µ. The number of microstates W (p) all corresponding to the macrostate = (2p − N ) of magnetization M µ, in which p magnetic moments µ are in a given direction and N − p in the opposite direction, is the number of choices of the p moments to be reversed among the total of N . It is thus equal to N! p p . Since N is macroscopic and in general p too, CN = is a very CN p!(N − p)! large number. The scope of this book is the analysis of systems in equilibrium ∗ : this means that one has waited long enough so that macroscopic physical parameters now keep the same values in time within very small ﬂuctuations (the macrostate is ﬁxed), whereas microscopic quantities may continue their evolution. Besides, even when the system is not isolated ∗ from its surroundings, if it has reached equilibrium, by deﬁnition there is no algebraical transport out of the system of matter, energy, etc. The approach toward equilibrium is achieved through interactions, bringing very small amounts of energy into play, yet essential to thermalization : in a solid, for example, electrons reach an equilibrium, among themselves and with the ions, through the collisions they suﬀer, for example, on ions vibrating at non-vanishing temperature, or impurities. Here the characteristic time is very short, of the order of 10−14 sec in copper at room temperature. For other processes this time may be much longer : the time for a drink to reach its equilibrium temperature in a refrigerator is of the order of an hour. If the system has no measurable evolution during the experiment, it is considered to be in equilibrium.

1.3.2

Time Average and Ensemble Average

Once equilibrium is reached, the time average of a physical parameter, which is likely to ﬂuctuate in a macrostate, is calculated using the classical probability density (1.7) or the quantum density operator (1.11), that is, for example

1 The

words marked by an asterisk are deﬁned in the glossary preceding Chapter 1.

12

Chapter 1. Statistical Description of Large Systems

1 A = T

t +T

dt t

A(ri , p i , t)D(ri , p i , t)dτ,

(1.13)

Would the same result be obtained by calculating the average at a given time on an assembly of macrostates prepared in a similar way ? A special case, allowing this issue to be understood, is that of an isolated system, of constant total energy E. In the N -particle phase space, the condition of constant energy deﬁnes a 6N − 1-dimension surface. The points representing the various systems prepared at the same energy E are on this surface. Is the whole surface described ? What is the time required for that ? This equivalence of the time average with the average on an ensemble of systems prepared in the same way is the object of the ergodic theorem, almost always applicable and that we will assume to be valid in all the cases treated in this course. In Statistical Physics, the common practice since Josiah Willard Gibbs (18391903) is indeed to take the average, at a ﬁxed time, on an assembly (an ensemble) of systems of the same nature, prepared under the same macroscopic conditions : this is the so-called ensemble average. As an example, consider a system, replicated a large number of times, which is exchanging energy with a reservoir. The energy distribution between the system and the reservoir differs according to the replica whereas, as will be shown in § 2.3.1, the system temperature is always that of the heat reservoir. Let N be the number of prepared identical systems, where N is a very large number. We consider a physical parameter f that can ﬂuctuate, in the present case the energy of each system. Among the N systems, this parameter takes the value fl in Nl of them ; consequently, the ensemble average of the parameter f will be equal to 1 fN = Nl fl (1.14) N l

When N tends to inﬁnity, Nl /N tends to the probability pl to achieve the value fl and the ensemble average tends to the average value calculated using these probabilities : lim fN = f = pl f l (1.15) N →∞

1.3.3

l

Equiprobability

On daily experience ( !), when throwing a die, in the absence of any other information, it is commonly assumed that every face is equally likely to occur, so that the probability of occurrence of a given face (1 for example) is 1/6.

Statistical Hypotheses ; Equiprobability

13

In Statistical Physics, an analogous postulate is stated : one assumes that in the absence of additional information, for an isolated system of energy ranging between E and E + δE, where δE is the uncertainty, all the accessible microstates are equally probable. (In a classical description, this is equivalent to saying that all the accessible phase space cells are equiprobable.) Considering the example of the magnetic moments, this means that, to obtain the same magnetization value, it is equally likely to have a localized magnetic moment reversed on a particular site or on another one. The larger the number of accessible states, the larger the uncertainty or disorder, the smaller the information about the system : a measure of the disorder is the number of accessible states, or the size of the accessible volume, of the phase space.

Ω/2

Ω/2

Fig. 1.4 : Joule-Gay-Lussac expansion. Consider for example a container of volume Ω, split into two chambers of volume Ω/2 each, isolated from its surroundings (Fig. 1.4). If the gas at equilibrium is completely located inside the left chamber, the disorder is smaller than if, at the same conditions, the equilibrium occurs in the total volume Ω. Opening the partition between the two chambers, in the so-called JouleGay-Lussac free expansion, means relaxing the constraint on the molecules to be located in the left chamber. This is an irreversible ∗ process2 as the probability for a spontaneous return, at a later time, to the situation where the N molecules all lie in the left chamber is extremely low (2−N ). It appears that a macroscopic initial state distinct from equilibrium will most probably evolve toward a larger disorder, although the ﬂuctuations in equilibrium are identical when the direction of time is reversed. The direction of time deﬁned in Statistical Physics is based on the idea that a statistically improbable state was built through an external action, requiring the expending of energy. The basic postulate of Statistical Physics, stated by Ludwig Boltzmann in 1877 for an isolated system, consists in deﬁning a macroscopic quantity from the number of accessible microstates. This is the statistical entropy S, given 2 A very interesting discussion on the character of the time direction and on irreversibility in Statistical Physics can be found in the paper by V. Ambegaokar and A.A. Clerk, American Journal of Physics, vol 67, p. 1068 (1999).

14

Chapter 1. Statistical Description of Large Systems

by S = kB ln W (E)

(1.16)

Here W (E) is the number of microstates with an energy value between E and E+δE ; the constant kB is a priori arbitrary. To make this statistical deﬁnition of the entropy coincide with its thermodynamical expression (see § 3.3), one should take kB equal to the Boltzmann constant, which is the ratio of the ideal gas constant R = 8.31 J/K to the Avogadro number N = 6.02 × 1023, that is, kB = R/N = 1.38 × 10−23J/K. This is what will be taken in this course. Take the example of the localized magnetic moments of a solid in the presence of an : in the conﬁguration where p moments are parallel external magnetic ﬁeld B and N − p moments are in the opposite direction, the magnetization is to B = (2p − N ) ·B = (N − 2p)µ · B. It was M µ and the magnetic energy E = −M N! . For a large system, shown in § 1.3.1 that in such a case W (E) = p!(N − p)! using the Stirling formula for the factorial expansion, one gets ïÅ ã Å ã E 2 kB N S = kB ln W (E) = 1+ ln 2 N µB B 1 + (E/N µB B) ã Å ãò (1.17) Å 2 E ln + 1− N µB B 1 − (E/N µB B) When two independent, or weakly coupled, subsystems are associated, with respective numbers of microstates W1 and W2 , the number of microstates of the combined system is W = W1 · W2 , since any microstate of subsystem 1 can be associated with any microstate of 2. Consequently, S = S1 + S2 , so that S is additive ∗ . A special case is obtained when the two subsystems are identical, then S = 2S1 , that is, S is extensive ∗ . Formula (1.16) is fundamental in that it bridges the microscopic world (W ) and the macroscopic one (S). It will be shown in § 3.3 that S is indeed identical to the entropy deﬁned in Thermodynamics. Before stating the general properties of the statistical entropy S as deﬁned by (1.16), we ﬁrst discuss those of W (E). The aim is to enumerate the microstates of given energy E, to the measurement accuracy δE. The number of accessible states is related to both the number of particles and to the number of degrees of freedom of each particle. It is equal to the accessible volume of the phase space, to the factor CN /h3N (§ 1.2.2). The special case of free particles, in motion in a macroscopic volume, is treated in Appendix 1.3. This example of macroscopic system allows one to understand : i) the extremely fast variation of W (E) with E, as a power of the particle number N , assumed to be very large (a similar fast variation is also obtained for an assembly of N particles subjected to a potential) ; ii) the to-

General Properties of the Statistical Entropy

15

tally negligible eﬀect of δE in practical cases ; iii) the eﬀective proportionality of ln W (E), and thus of S, to N . In the general case, the particles may have other degrees of freedom (in the example of diatomic molecules : rotation, vibration), they may be subjected to a potential energy, etc. Results i), ii), iii) are still valid.

1.4

General Properties of the Statistical Entropy

As just explained on a special case, the physical parameter that contains the statistical character of the system under study is the statistical entropy. We now deﬁne it on more general grounds and relate it to the information on the system.

1.4.1

The Boltzmann Deﬁnition

The expression (1.16) of S given by Boltzmann relies on the postulate of equal probability of the W accessible microstates of an isolated system. S increases as the logarithm of the microstates number.

1.4.2

The Gibbs Deﬁnition

A more general deﬁnition of the statistical entropy was proposed by Gibbs : S = −kB

pi ln pi

(1.18)

i

This deﬁnition applies to an assembly of systems constructed under the same macroscopic conditions, made up in a similar way according to the process deﬁned in § 1.3.2 ; in the ensemble of all the reproduced systems, pi is the probability to attain the particular state i of the system, that is, the microstate i. It is this Gibbs deﬁnition of entropy that will be used in the remaining part of this course. For example, for systems having reached thermal equilibrium at temperature T , we will see that the probability of realizing a microstate i is higher, the lower the microstate energy (pi is proportional to the Boltzmann factor, introduced in § 2.4.1). The Boltzmann expression (1.16) of the entropy is a special case, which corresponds to the situation where all W microstates are equiprobable, so that the probability of realizing each of them is pi = 1/W .

16

Chapter 1. Statistical Description of Large Systems

Then : S = −kB ln

1 ( pi ) = kB ln W W i

(1.19)

It is not possible to proceed the other way and to deduce the general Gibbs deﬁnition from the particular Boltzmann deﬁnition. Yet let us try to relate the Gibbs and the Boltzmann deﬁnitions using a particular example : we are considering a system made up of n1 + n2 equiprobable microstates regrouped into two subsystems 1 and 2. Subsystem 1 contains n1 microstates, its probability of realization is p1 = n1 /(n1 + n2 ). In the same way, subsystem 2 contains n2 microstates, its probability of realization is equal to p2 = n2 /(n1 + n2 ). From the Boltzmann formulation, the statistical entropy of the total system is Stot = kB ln(n1 + n2 )

(1.20)

This entropy is related to the absence of information on the microstate of the total system in fact realized : – we do not know whether the considered microstate belongs subsystem 1 or 2 : this corresponds to the entropy S12 , which is the Gibbs term that we want to ﬁnd in this example ; – a part S1 of the entropy expresses the uncertainty inside system 1 owing to its division into n1 microstates S1 = p1 (kB ln n1 )

(1.21)

and is weighted by the probability p1 . In the same way S2 comes from the uncertainty inside system 2 S2 = p2 (kB ln n2 )

(1.22)

Since the various probabilities of realization are multiplicative, the corresponding entropies, which are related to the probabilities through logarithms, are additive : Stot = S1 + S2 + S12 kB ln(n1 + n2 ) =

n1 n2 kB ln n1 + kB ln n2 + S12 n1 + n2 n1 + n2

(1.23)

General Properties of the Statistical Entropy

17

Then the S12 term is given by Å S12 = kB Å

n1 + n2 n1 + n2

ã ln(n1 + n2 )

− kB

ã Å ã n1 n2 ln n1 − kB ln n2 n1 + n2 n1 + n2

= −kB

n1 n2 n1 n2 ln − kB ln n1 + n2 n1 + n2 n1 + n2 n1 + n2

(1.24)

= −kB (p1 ln p1 + p2 ln p2 ) Indeed this expression is the application of the Gibbs formula to this particular case. On the other hand, expression (1.16) of the Boltzmann entropy is a special case of SGibbs , which makes (1.18) maximum : the information on the system 1 for each is the smallest when all W microstates are equiprobable and pi = W i. The Boltzmann expression is the one that brings S to its maximum with the constraint of the ﬁxed value of the system energy. The expression (1.18) of S, as proposed by Gibbs, satisﬁes the following properties : S ≥ 0, for 0 ≤ pi ≤ 1 and ln pi ≤ 0 S = 0 if the system state is certain : then p = 1 for this particular state, 0 for the other ones. S is maximum when all pi ’s are equal. S is additive : indeed, when two systems 1 and 2 are weakly coupled, so that ⎧ ⎪ ⎪ ⎨ S1

= −kB

⎪ S ⎪ ⎩ 2

= −kB

i

p1i ln p1i p2j ln p2j

(1.25)

j

by deﬁnition, the entropy of the coupled system will be written : Stot = −kB

pl ln pl

(1.26)

l

Now, since the coupling is weak, each state of the total system is characterized

18

Chapter 1. Statistical Description of Large Systems

by two particular states of subsystems 1 and 2 and pl = p1i p2j . Consequently, Stot = −kB p1i p2j (ln p1i + ln p2j ) i,j

= −kB

i

p1i ln p1i

p2j + p2j ln p2j p1i

j

j

(1.27)

i

= S1 + S 2 One thus ﬁnds the additivity of S, its extensivity being a special case that is obtained by combining two systems of the same density, but of diﬀerent volumes.

1.4.3

The Shannon Deﬁnition of Information

We just saw that the more uncertain the system state, or the larger the number of accessible microstates, the larger S ; in addition, S is zero if the system state is exactly known. The value of S is thus related to the lack of information on the system state. The quantitative deﬁnition of information, given by Claude Shannon (1949), is closely copied from the Gibbs deﬁnition of entropy.3 It analyzes the capacity of communication channels. One assumes that there are W possible distinct messages, that si (i = 0, 1, . . . W − 1) is the content of the message number i, and that the probability for si to be emitted is pi . Then the information content I per sent message is I=−

W −1

pi log2 pi

(1.28)

i=0

This deﬁnition is, to the constant kB ln2, the same as that of the Gibbs entropy, since log2 pi = lnpi /ln2. The entropy is also related to the lack of information : probabilities are attributed to the various microstates, since we do not exactly know in which of them is the studied system. These probabilities should be chosen in such a way that they do not include unjustiﬁed hypotheses, that is, only the known properties of the system are introduced, and the entropy (the missing information) is maximized with the constraints imposed by the physical conditions of the studied system. This is the approach that will be chosen in the arguments of § 2.3, 2.4, and 2.5. 3 See the paper by J. Machta in American Journal of Physics, vol 67, p.1074 (1999), which compares entropy, information, and algorithmics using the example of meteorological data.

Summary of Chapter 1 The time evolution of an individual particle is associated with a trajectory in the six-dimensional phase space, of coordinates (r, p ). The Heisenberg uncertainty principle and the classical limit of quantum properties require that a quantum state occupies a cell of area h3 in this space. At a given time t an N -particle state is represented by a point in the 6N -dimensional phase space. The classical statistical description of a macroscopic system utilizes a probability density, such that the probability of ﬁnding the system of N particles N

1 , . . . , pN ), to Π d3ri d3 pi , is in the neighborhood of the point (r1 , . . . , rN , p i=1

equal to N CN 3 1 , . . . , pN ) 3N d ri d pi D(r1 , . . . , rN , p h i=1

where CN is a constant depending on N only. In Quantum Mechanics the density operator ˆ = D pn |ψn ψn | n

is introduced, which contains the uncertainties related both to the incomplete knowledge of the system and to the quantum measurement. A conﬁguration deﬁned by the data of the microscopic physical parameters is a microstate. A macrostate is deﬁned by the value of macroscopic physical parameters ; it is generally produced by a very large number of microstates. This course will only deal with Statistical Physics in equilibrium. The time average of a ﬂuctuating physical parameter is generally equivalent to the average on an assembly of identical systems prepared in the same way (“ensemble average”). One assumes that, in the absence of additional information, in the case of an 19

20

Summary of Chapter 1

isolated system of energy between E and E + δE, where δE is the uncertainty, all the W (E) accessible microstates are equally probable. The statistical entropy S is then deﬁned by S = kB ln W (E) where kB = R/N is the Boltzmann constant, with R the ideal gas constant and N the Avogadro number. The general deﬁnition of the statistical entropy, valid for an ensemble average, is S = −kB pi ln pi i

where pi is the probability of occurrence of the particular microstate i of the system. The expression for an isolated system is a special case of the latter deﬁnition. The statistical entropy is an extensive parameter, which is equal to zero when the microscopic state of the system is perfectly known.

Appendix 1.1 The Liouville Theorem in Classical Mechanics There are two possible ways to consider the problem of the evolution between the times t and t + dt of an ensemble of systems taking part in the ensemble average, each of them containing N particles, associated with the same macrostate and prepared in the same way. To one system, in its particular microstate, corresponds one point in the 6N -dimensional phase space. Either we may be concerned with the time evolution of systems contained in a small volume around the considered point (§ A) ; or we may evaluate the number of systems that enter or leave a ﬁxed inﬁnitesimal volume of the phase space during a time interval (§ B). In both approaches it will be shown that the elementary volume of the phase space is conserved between the times t and t + dt. A. Time evolution of an ensemble of systems At time t the systems occupy the phase space volume dτ around the point with coordinates (q1 , . . . qi , . . . , q3N , p1 , . . . pi , . . . , p3N ) (Fig. 1.5). The inﬁnitesimal 3N volume has the expression dτ = CN dpi dqi , where CN is a constant that i=1

only depends on N . To compare this volume to the one at the instant t + dt, one has to evaluate 3N 3N the products dpi (t)dqi (t) and dpi (t + dt)dqi (t + dt), that is, to calculate i=1

i=1

the Jacobian of the transformation ⎧ ⎪ ⎪ pi (t) → pi (t + dt) = pi (t) + dt · ∂pi (t) ⎨ ∂t ⎪ ∂q ⎪ ⎩ qi (t) → qi (t + dt) = qi (t) + dt · i (t) ∂t 21

(1.29)

22

Appendix 1.1

pi t + dt dτ t

dτ O

qi

Fig. 1.5: The diﬀerent systems, which were located in the volume dτ of the 6N -dimensional phase space at t, are in dτ at t + dt. i.e., the determinant of the 6N × 6N matrix ⎤ ⎡ ∂qi (t + dt) ∂pi (t + dt) ⎢ ∂qj (t) ∂qj (t) ⎥ ⎥ ⎢ ⎥ ⎢ ⎣ ∂qi (t + dt) ∂pi (t + dt) ⎦ ∂pj (t) ∂pj (t)

(1.30)

The Hamilton equations are used, which relate the conjugated variables qi and pi through the N -particle hamiltonian and are the generalization of Eq. (1.4) : ⎧ ∂qi ∂H ⎪ ⎪ ⎨ ∂t = ∂p i (1.31) ⎪ ∂pi ∂H ⎪ ⎩ =− ∂t ∂qi One deduces

and in particular

⎧ ∂H ⎪ ⎪ ⎨ qi (t + dt) = qi (t) + dt · ∂p i ⎪ ∂H ⎪ ⎩ pi (t + dt) = pi (t) − dt · ∂qi

(1.32)

⎧ ∂2H ∂qi (t + dt) ⎪ ⎪ ⎨ ∂q (t) = δij + dt · ∂q ∂p j j i 2 ⎪ ∂p (t + dt) ∂ H ⎪ ⎩ i = δij − dt · ∂pj (t) ∂pj ∂qi

(1.33)

Appendix 1.1

23

The obtained determinant is developed to ﬁrst order in dt : ˆ ) = 1 + dt · Tr M ˆ + O(dt2 ) Det(1 + dt · M with ˆ =− Tr M

3N 3N ∂2H ∂2H + =0 ∂pi ∂qi i=1 ∂pi ∂qi i=1

(1.34)

Consequently, the volume around a point of the 6N -dimensional phase space is conserved during the time evolution of this point ; the number of systems in this volume, i.e., D(. . . qi , . . . , . . . pi , . . . , t)dτ , is also conserved. B. Number of systems versus time in a ﬁxed volume dτ of the phase space In the volume dτ = CN

3N

dqi · dpi , at time t, the number of systems present

i=1

is equal to D(q1 , . . . qi , . . . q3N , p1 , . . . pi , . . . p3N )dτ

(1.35)

Each system evolves over time, so that it may leave the considered volume while other systems may enter it. p1

q˙1 (t)dt dq1 dp1 (q˙1 (t) + O

∂ q˙1 (t) dq1 ) dt ∂q1 q1

Fig. 1.6: In the volume dq1 · dp1 (thick line), between the times t and t + dt, the systems which enter were in the light grey rectangle at t ; those which leave the volume were in the dark grey hatched rectangle. It is assumed that q˙1 (t) is directed toward the right. Take a “box”, like the one in Fig. 1.6 which represents the situation for the only coordinates q1 , p1 . If only this couple of conjugated variables is considered, the

24

Appendix 1.1

problem involves a single coordinate q1 in real space (so that the velocity q˙1 is necessarily normal to the “faces” at constant q1 ) and we determine the variation of the number of systems on the segment for the q1 coordinate [q1 , q1 + dq1 ]. Considering this space coordinate, between the times t and t + dt, one calculates the diﬀerence between the number of systems which enter the volume through the faces normal to dq1 and the number which leave through the same faces : the ones which enter were at time t at a distance of this face smaller or equal to q˙1 (t)dt (light grey area on the Fig.), their density is D(q1 − dq1 , . . . qi , . . . qÅ 3N , p1 , . . . pi , . . . p3Nã). Those which leave were inside the ∂ q˙1 (t) dq1 dt from the face (dark grey hatched volume, at a distance q˙1 (t) + ∂q1 area). The resulting change, between t and t + dt and for this set of faces, of the number of particles inside the considered volume, is equal to −

∂(q˙1 (t)D) dq1 dt ∂q1

(1.36)

For the couple of “faces” at constant p1 , an analogous evaluation is performed, also contributing to the variation of the number of particles in the considered “box.” Now at 3N dimensions, for all the “faces,” the net increase of the number of particles in dτ during the time interval dt is −

3N ∂(Dq˙i (t)) i=1

∂qi

dqi dt −

3N ∂(Dp˙ i (t)) i=1

∂pi

dpi dt

(1.37)

This is related to a change in the particle density in the considered volume ∂D through to dtdτ . Thus the net rate of change is given by ∂t ã ã Å 3N Å 3N ∂D ∂D ∂D ∂ q˙i ∂ p˙ i (1.38) =− q˙i + p˙ i − D + ∂t ∂qi ∂pi ∂qi ∂pi i=1 i=1 From the Hamilton equations, the factor of D is zero and the above equation is equivalent to ã 3N Å dD ∂D ∂D ∂D = + q˙i + p˙ i = 0 dt ∂t ∂qi ∂pi i=1

(1.39)

dD(q1 , . . . , qi , . . . q3N , p1 , . . . , pi , . . . p3N ) is the time derivative of the where dt probability density when one moves with the representative point of the system in the 6N -dimensional phase space. This result is equivalent to the one of § A.

Appendix 1.1

25

As an example, consider the microcanonical ensemble, in which the density D is a constant for the energies between E and E + δE, zero elsewhere. The density D is function of the sole energy, a constant in the microcanonical case, so that its derivative with respect to energy is zero between E and E + δE. Then D is independent of the coordinates qi and of the momenta pi , and from (1.39) it is time-independent.

Appendix 1.2 Properties of the Density Operator in Quantum Mechanics ˆ deﬁned by The operator D ˆ = D

pn |ψn ψn |

n

where pn is the probability to realize the microstate |ψn , has the following properties : ˆ =D ˆ+ i) it is hermitian : D Indeed, for any operator |u, ˆ u|D|u =

ˆ + |u pn |u|ψn |2 = u|D

(1.40)

n

ii) it is deﬁned positive ˆ u|D|u =

pn |u|ψn |2 ≥ 0

(1.41)

n

iii) it is normed to unity, since ˆ= Tr D

n

27

pn = 1

(1.42)

Appendix 1.3 Estimation of the Number of Microstates for Free Particles Conﬁned Within a Volume Ω Take the example of a free point particle, of energy and momentum linked p2 = ε ; in three dimensions, the particle occupies the volume Ω. The by 2m constant energy surfaces correspond to a constant p. In the 6-dimensional (r, p ) phase space for a single particle, the volume associated with the energies between ε and ε + δε is equal to : Ω4πp2 δp = Ω4πp.pδp, with … m pδp , that is, δp = δε (1.43) δε = m 2ε For a single particle, the number of accessible states between ε and ε + δε is proportional to Ωp · pδp, i.e., Ωε1/2 δε, i.e., also to Ωε(3/2)−1 δε. For N independent free particles, the constant-energy sphere of the 6N dimensional phase space contains all the states of energy smaller or equal to E, where E is of the order of N ε. The volume of such a sphere varies in ΩN p3N . The searched microstates, of energy between E and E + δE, correspond to a volume of the order of ΩN p3N −1 δp, that is, ΩN p3N −2 δE, for pδp ∝ δE. The number W (E) of microstates with energy between E and E + δE is proportional to this volume, i.e., Å W (E) =

Ω Ω0

ãN Å

E E0

ã3N/2

δE E

(1.44)

where the constants Ω0 and E0 provide the homogeneity of the expression and where δE, the uncertainty on the energy E, is very much smaller than this energy value. The increase of W (E) versus energy and volume is extremely fast : for 29

30

Appendix 1.3

example, if N is of the order of the Avogadro number, N , Å W (E) =

Ω Ω0

ã6×1023 Å

E E0

ã9×1023

δE E

(1.45)

which is such a huge number that it is diﬃcult to imagine it ! The logarithm in the deﬁnition of the Boltzmann entropy S is here : ã Å Ω E 3N δE ln W (E) = N ln ln (1.46) + + ln Ω0 2 E0 E E is of the order of N times the energy of an individual particle. For N in the N range, the dominant terms in ln W (E) are proportional to N : indeed for an experimental accuracy δE/E of the order of 10−4 , the last logarithm is equal to −9.2 and is totally negligible with respect to the two ﬁrst terms, of the order of 1023 . Note : the same type of calculation, which estimates the number of accessible states in a given energy range, will also appear when calculating the density of states for a free particle (§ 6.4).

Chapter 2

The Diﬀerent Statistical Ensembles. General Methods in Statistical Physics In the ﬁrst chapter we convinced ourselves that the macroscopic properties of a physical system can only be analyzed through a statistical approach, in the framework of Statistical Physics. Practically, the statistical description of a system with a very large number of particles in equilibrium, and the treatment of any exercise or problem of Statistical Physics, always follow the two steps (see the section General method for solving exercises and problems, at the end of Ch. 9) : – First, determination of the microscopic states (microstates) accessible by the N -particles system : this is a Quantum Mechanics problem, which may be reduced to a Classical Mechanics one in some speciﬁc cases. § 2.1 will schematically present the most frequent situations. – Second, evaluation of the probability for the system to be in a particular microstate, in the speciﬁed physical conditions : it is at this point that the statistical description comes in. To solve this type of problem, one has to express that in equilibrium the statistical entropy is maximum, under the constraints deﬁned by the physical situation under study : the system may be isolated, or in thermal contact 31

32

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

with a heat reservoir (a very large system, which will dictate its temperature), the system may exchange particles or volume with another system, and so forth. The experimental conditions deﬁne the constraint(s) from the conservation laws adapted to the physical situation (conserved total energy, conserved number of particles, etc.). In this chapter, the most usual statistical ensembles will be presented : “microcanonical” ensemble in § 2.2, “canonical” ensemble in § 2.4, “grand canonical” ensemble in § 2.5, using the names introduced at the end of the 19th century. In all these statistical ensembles, one reproduces the considered system in a thought experiment, thus realizing an assembly of macroscopic objects built under the same initial conditions, which allows average values of ﬂuctuating physical parameters to be deﬁned. For a given physical condition, one will decide to use the statistical ensemble most adapted to the analysis of the problem. However, as will be shown, for systems with a macroscopic number of particles (the only situation considered in this book), there is equivalence, to a extremely small relative ﬂuctuation, between the various statistical ensembles that will be described here : in the case of a macroscopic system, a measurement will not distinguish between a situation in which the energy is ﬁxed and another one in which the temperature is given. It is only the convenience of treatment of the problem that will lead us to choose one statistical ensemble rather than another. (On the contrary for systems with a small number of particles, or of size intermediate between the microscopic and macroscopic ranges, the socalled “mesoscopic systems,” nowadays much studied in Condensed Matter Physics, the equivalence between statistical ensembles is not always valid.)

2.1

Determination of the Energy States of an N -Particle System

For an individual particle i in translation, of momentum pi and mass m, under a potential energy Vi (ri ), the time-independent Schroedinger equation (eigenvalues equation) is Ç ˆ i |ψ αi = h i =

å pˆi 2 + Vi (ri ) |ψiαi 2m

(2.1)

αi i εα i |ψi

where αi expresses the diﬀerent states accessible to this particle. As soon as several particles are present, one has to consider their interactions, which leads to a several-body potential energy. Thus, the hamiltonian for N particles in motion, under a potential energy and mutual interactions, is given

Energy States of an N -Particle System

33

by ˆ = H

pˆi 2 i

2m

+ Vi (ri ) +

Vij (ri , rj )

(2.2)

i<j

This is the case, for example, for the molecules of a gas, conﬁned within a limited volume and mutually interacting through Van der Waals interactions ; this is also the situation of electrons in solids in attractive Coulombic interaction with the positively charged nuclei (Vi term), and in repulsion between themselves (Vij term). The latter interactions, expressed in the Vij ’s, are essential in the approach toward thermal equilibrium through the transitions they induce between microscopic states, but they complicate the solution of the N -particle eigenstate problem. Now they mostly correspond to energy terms very small with respect to those associated with the remaining part of the hamiltonian. An analogous problem is faced when one is concerned by the magnetic moments of a ferromagnetic material : the hamiltonian of an individual electronic intrinsic magnetic moment located in a magnetic ﬁeld is given by ˆ = − h µˆB · B

(2.3)

ˆ by : where the Bohr magneton operator µˆB is related to the spin operator S e ˆ µ ˆB = − S m

(2.4)

and has for eigenvalues ∓e/2m (see a course of Quantum Mechanics). The ha contains, miltonian of an ensemble of magnetic moments in a magnetic ﬁeld B ˆ other terms expressing their mutual interin addition to terms similar to h, actions, i.e. ˆ =− + µˆBi · B ˆBi · µˆBj H Jij µ (2.5) i

i<j

where Jij is the coupling term between the moments localized at sites i and j. According to the site i or j, the moment orientation is diﬀerent. All these types of hamiltonians, in which several-particle terms appear, are generally treated through approximations, which allow them to be reduced to ˆ is written as a sum of similar hamiltonians, the simplest situation where H each concerning a single particle. There are several methods to achieve such a reduction : – either neglecting the interactions between particles in the equilibrium state, as they correspond to very small energies. As a consequence, in the above example of the gas molecules, the only term included in the potential energy

34

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

expresses the conﬁnement with a box (conﬁnement potential) or an external ﬁeld (gravity) ; the particles are then considered as “free” ; – or using the so-called “mean ﬁeld” treatment : each particle is subjected to a mean eﬀect from the other ones. This will be a repulsive potential energy in the case of the Coulombic interaction between electrons in a solid, an eﬀective magnetic ﬁeld equivalent to the interactions from the other magnetic moments in the case of a ferromagnetic material ; ˆ in order to decouple the – or changing of variable in the hamiltonian H, ˆ variables, so that H may be re-written as a sum of terms (“normal modes” deﬁning quasi-particles, for example, in the problem of the atomic vibrations in a crystal at a given temperature). From now on, let us assume that the total-system hamiltonian is indeed expressed as the sum of similar individual-particle hamiltonians ˆ i ψ αi (ri ) = εαi ψ αi (ri ) ˆ i , with h ˆ = (2.6) h H i i i i

Then it will only be necessary to solve the problem for an individual particle : indeed you learnt in Quantum Mechanics that the solution of ˆ r1 , r2 , . . . , rN ) = Eψ(r1 , r2 , . . . , rN ) Hψ(

(2.7)

is

ˆj h

N

j

i=1

N α ψiαi (ri ) = εj j ψiαi (ri ) j

(2.8)

i=1

i.e., E=

i

i εα r1 , r2 , . . . , rN ) = i , ψ(

N

ψiαi (ri )

(2.9)

i=1

The total energy of the system is the sum of the individual particle energies and the N -particle eigenfunction is the product of the eigenfunctions for each single particle. In the ﬁrst three chapters of this book, the problems studied will concern discrete variables (for example„ spin magnetic moments) or particles distinguishable by their ﬁxed coordinates, like the atoms of a solid vibrating around their equilibrium position at ﬁnite temperature. The systems with translation degrees of freedom will be analyzed in the classical framework in chapter 4 (ideal gas), or in the framework of the Quantum Statistics from chapter 5 until the end of the course : in the Quantum Statistics description of the accessible

Isolated System in Equilibrium : “Microcanonical Ensemble”

35

microstates, one has to account for the indistinguishability of the particles, as expressed by the Pauli principle. In all that follows, we will assume that the microscopic states of the considered system are known. We are going to rely on the hypothesis of maximum entropy, as introduced in chapter 1, § 1.3 and § 1.4, under the speciﬁc physical constraints due to the given experimental conditions ; we will then deduce the system macroscopic properties, using Statistical Physics.

2.2

Isolated System in Equilibrium : “Microcanonical Ensemble”

In such a system one assumes that there is no possible exchange with its surroundings (Fig. 2.1) : the volume is ﬁxed, the number of particles N is given, the total energy assumes a ﬁxed value E, i.e., its lies in the range between E E + δE, where δE is the uncertainty. This situation is called “microcanonical ensemble.” It is the framework of application of the Boltzmann hypothesis of equiprobability of the W (E) accessible microstates (§ 1.3.3). The probability to realize any of these microstates is equal to 1/W (E) ; the statistical entropy S is maximum under these conditions (with respect to the situation where the probabilities would not be equal) and is given by S = kB ln W (E)

(2.10)

E to δE

Fig. 2.1 : Isolated system. Its energy has a ﬁxed value. To solve this type of problem, one has to determine W (E) : this means performing a combinatory calculation of the number of occurrences of the macroscopic energy E from the various conﬁgurations of the microscopic components of the system. For example, for an ensemble of localized magnetic moments located in an external magnetic ﬁeld, to each ﬁxed value of E corresponds a value of the macroscopic magnetization. The corresponding number of microstates W (E) was evaluated in § 1.3.1. The other physical macroscopic parameters

36

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

are deduced from S and its partial derivatives (see § 3.5). In particular a microcanonical temperature is deﬁned from the partial derivative of S versus E : 1/T = ∂S/∂E.

2.3

Equilibrium Conditions for Two Systems in Contact, Only Exchanging Energy

E A

A ∂S = kB β ∂E

∂S = kB β ∂E

Fig. 2.2: The systems A and A only exchange energy, the combined system A0 =A+A is isolated. Two distinct systems A and A are brought into thermal contact. The system A0 , made up of the combination of A and A (Fig. 2.2) is the isolated system, to which we can apply the above approach, i.e. look for the maximum of its statistical entropy. The systems in contact, A and A , can only exchange energy and are weakly coupled, which means that a possible interaction energy is neglected. We can also say that the system A under study has the constraint to be coupled to A , the total energy E + E being conserved. This is a very common situation, like the one of a drink can in a refrigerator, the ensemble (can-refrigerator) being taken as thermally insulated from the room in which the refrigerator stands, the room representing the surrondings.

2.3.1

Equilibrium Condition : Equal β Parameters

The total energy and the total statistical entropy are shared between A and A : E0 = E + E SA0 (E0 , E) = SA (E) + SA (E0 − E)

(2.11) (2.12)

From the Boltzmann relation (2.10), the equation (2.12) is equivalent to W0 (E0 , E) = WA (E) · WA (E0 − E)

(2.13)

Equilibrium Conditions for Two Systems in Contact

37

since for weakly coupled systems the number of microstates of the combined system is the product of the numbers of microstates of the two subsystems. After a long enough time has elapsed, the combined system (can-refrigerator) A0 reaches an equilibrium situation. Since the combination A0 of the two systems A and A is an isolated system, in equilibrium its entropy is maximum with respect to the exchange of energy between A and A : SA0 (E0 , E) = SA (E) + SA (E0 − E)

(2.14)

∂SA0 (E0 , E) ∂SA0 (E0 , E) ∂SA (E) ∂SA (E0 − E) = 0 , i.e., = + =0 ∂E ∂E ∂E ∂E (2.15) Once the equilibrium has been reached : ∂SA (E0 − E) ∂SA (E ) ∂SA (E) =− = (2.16) ∂E ∂E ∂E The obtained equilibrium condition is that the partial derivative of entropy versus energy takes the same value in both systems in contact, A and A , all the other parameters (volume, etc.) being kept constant. This derivative will be written : ∂S = kB β (2.17) ∂E where kB is the Boltzmann constant. ˜ one In equilibrium, reached for an energy value of system A equal to E = E, thus has β = β

(2.18)

We know that the entropy S and the energy E are both extensive* parameters : by deﬁnition the values of extensive parameters are doubled when the system volume Ω is doubled together with the number N of particles, while the density N/Ω, an intensive* parameter, is maintained in such a transformation. In such a process, the parameter β is not changed, it is an intensive parameter having the dimension of reciprocal energy, the dimension of kB β being a reciprocal temperature. In § 3.3 the partial derivative of the entropy versus energy will be identiﬁed with 1/T , where T is the absolute temperature, so that (2.18) expresses that in equilibrium T = T .

2.3.2

Fluctuations of the Energy Around its Most Likely Value

The number of microstates of the total system A0 realizing the splitting of the energies [E in A, E0 − E in A ] is proportional to the product

38

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

WA (E) · WA (E0 − E). The probability of occurrence of the corresponding microstate is WA (E)WA (E0 − E) WA (E)WA (E0 − E) = p(E, E0 − E) = Wtot (E0 ) WA (E)WA (E0 − E)

(2.19)

E

Let us consider the variation of this probability versus E around the value ˜ achieving the entropy maximum with respect to an energy exchange betE ween A and A . In (2.19), only the numerator is a function of E, since the denominator is a sum over all the possible values of E. For a macroscopic system, W (E) increases very fast with E, as already veriﬁed on the example of free particles in Appendix 3 of Chapter 1. On the other hand, WA (E0 − E) strongly decreases with E if A is macroscopic. Their product goes through a very steep maximum, corresponding to the ˜ state of maximum probability of the combined system, reached for E = E. This energy is very close to the average value E of the energy in system A, ˜ as the probability practically vanishes outside the immediate vicinity of E. ˜ The probability variation around E is much faster than that of its logarithm, which varies versus E in the same way as SA0 (E) = kB ln[WA (E)WA (E0 − E)]

(2.20)

because the combined system A0 is isolated. Note that in the particular case of Å a free-particle system, where W (E) ∼ E N ã ∂S (see Appendix 3, Chapter 1), β = is of the order of N/E. Since the ∂E Ω ˜ the average energy per particle is of the order total energy is of the order E, 1/β = kB T , to a numerical factor of order unity. The average energy of a free particle at temperature T is thus of order kB T . Now we look for the probability variation versus the energy of system A ˜ The logarithm of this probability is for E close to the most likely value E. proportional to the entropy of the combined system SA0 (E), considered as a function of E, as A0 is an isolated system to which the Boltzmann formula SA0 (E) = kB ln W0 (E) applies. In equilibrium, the entropy of the combined system A0 is maximum with respect to an energy exchange between the two systems ; the entropy of A0 ˜ small : can be developed around this maximum, for E − E ã Å 2 Å ã ˜ + ∂SA0 ˜ + 1 ∂ S A0 ˜ 2 (2.21) (E − E) (E − E) SA0 (E) = SA0 (E) ∂E E˜ 2 ∂E 2 E˜

˜ 3 + O (E − E)

Equilibrium Conditions for Two Systems in Contact

39

Here the ﬁrst derivative is equal to zero and the second derivative is negative, as the extremum is a maximum. This second derivative includes the contributions of both A and A and is equal to : ∂ 2 SA ∂ 2 S A ∂ 2 S A0 = + = kB 2 2 ∂E ∂E ∂E 2

Å

∂β ∂β + ∂E ∂E

ã = −kB λ0

(2.22)

˜ In fact, since β = 1/kB T , staAll these terms are calculated for E = E. ting that ∂β/∂E is negative means that the internal energy increases with temperature : this is always realized in practical cases.1 . From the entropy properties (2.21) and (2.22), using the Boltzmann relation one deduces those of the number of accessible microstates : ˜ exp − λ0 (E − E) ˜ 2 /2 W0 (E) ≈ W0 (E)

(2.23)

The probability of realizing the state with the sharing of the energies [E in system A and E0 − E in system A ] is proportional to W0 (E). It is given by

˜ E0 − E) ˜ exp − λ0 (E − E) ˜ 2 /2 p(E, E0 − E) ≈ p(E,

(2.24)

˜ This √ is a Gaussian function around the equilibrium state E. Its width ∆E = 1/ λ0 depends of the size of the total system like Å

∂β ∂β + ∂E ∂E

ã−1/2 (2.25)

that is, like the square root of a number of particles, since E and E are ˜ extensive and the parameters β and β intensive. The relative width ∆E/E varies as a reciprocal number of particles. For a macroscopic system, where N and N ∼ 1023 , this relative width is of the order of 10−11 , thus unmeasurable : the ﬂuctuations of the energy of system A around its equilibrium value are practically null. This is equivalent to saying that, for a system with a very large number of particles, situations i), in which the energy of A is exactly ˜ and ii), in which system A reached an equilibrium equal to the value E, ˜ energy E through exchanges with a heat reservoir, lead to the same physical measurement of the energy. 1 When the total energy has an upper bound, as in the case of the spins in a paramagnetic solid, the maximum energy is reached when all the magnetic moments are antiparallel to the applied magnetic ﬁeld. In the vicinity of this maximum, adding energy produces a decrease in entropy, so that (2.17) leads to a negative β, thus to a “ negative temperature.” However in the situation of a real paramagnetic solid, the spin degree of freedom is in equilibrium with the other degrees of freedom, like the ones related to the atoms’ vibrations around their equilibrium positions. For the latter ones, the energy is not limited and for the whole solid, the total energy indeed increases with temperature.

40

2.4

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

System in Contact with a Heat Reservoir, “Canonical Ensemble”

When two systems are brought into thermal contact, one being much larger than the other (the can is much smaller than the refrigerator) (Fig. 2.3), the larger system behaves like a heat reservoir of energy E0 − E close to E0 , the energy E of the other system thus being very small. The statistical entropy, and, consequently, the parameter β of the larger system, are close to those of the total system : Å Å 2 ã ã ã Å ∂SA (E) ∂ SA (E) ∂SA = − E + O(E 2 ) (2.26) ∂E E0 −E ∂E ∂E 2 E0 E0 Å 2 ã ∂ SA (E) + O(E 2 ) (2.27) = kB β0 − E ∂E 2 E0

E A

A β

Fig. 2.3: The system A is in contact with the heat reservoir A which dictates its parameter β to A.

When the equilibrium is reached, the parameter β of the smaller system is adjusted to the parameter β of the larger system, which itself takes a value close to that of the parameter β0 of the ensemble : the larger system acts as a heat reservoir or thermostat, a body that keeps its value of β and dictates this parameter to the smaller system (we will see in chapter 3 that this is equivalent to saying that it keeps its temperature when put in thermal contact : for example, in equilibrium the can takes the temperature of the refrigerator). The smaller system A can be a macroscopic system, suﬃciently smaller than A , or even a single microscopic particle, like an individual localized magnetic moment inside a paramagnetic solid, with respect to which the rest of the macroscopic solid plays the role of a reservoir. The situation in which the system under study is in thermal contact with a heat reservoir that gives its temperature T to the system is called “canonical,” since it is a very frequent situation among the studied problems. The ensemble

Canonical Ensemble

41

of systems in contact with the heat reservoir, on which we will calculate the ensemble average, is the “canonical ensemble.”

2.4.1

The Boltzmann Factor

We are thus in the situation where a smaller system A is coupled to a larger system A that dictates its value of the parameter β, which is close to the value for the combined isolated system. First consider the probability pi that a speciﬁc microstate i of system A, of energy Ei , is produced. It only depends on the properties of the heat reservoir and is the ratio of the number of microstates Wres (E0 − Ei ) of the reservoir with this energy to the sum of all the numbers of microstates for all the reservoir energies E0 − Ei : Wres (E0 − Ei ) pi = Wres (E0 − Ei )

(2.28)

i

Wres (E0 −Ei ) is calculated in the microcanonical ensemble for the given energy E0 − Ei and is related to Sres (E0 − Ei ) through the Boltzmann relation ln Wres (E0 − Ei ) =

1 Sres (E0 − Ei ) kB

(2.29)

Indeed, once the reservoir energy E0 − Ei is ﬁxed, the diﬀerent microstates in this situation are equally likely. Since

Å Sres (E0 − Ei ) = SA (E0 ) − Ei

∂SA ∂E

ã + ...

(2.30)

E =E0

ln Wres (E0 − Ei ) = ln WA (E0 ) − Ei β0 + . . .

(2.31)

= constant − β0 Ei + . . . the probability of occurrence of this microstate is proportional to Wres (E0 − Ei ), i.e., pi = C exp(−β0 Ei )

(2.32)

In the probability pi there appears the exponential of the energy Ei of the system. This is the so-called “Boltzmann factor” (1871), or “canonical distribution,” an expression that will be used very often in this course, every time we will study an ensemble of distinguishable particles, in ﬁxed number 1 (the case of indistinguishable particles will N , of given temperature T = kB β

42

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

be treated in the framework of the Quantum Statistics in chapter 5 and the following ones). As an example, Fig. 2.4 schematizes the Boltzmann factor for a 4-microstate system. E

E β1 E4

E3

E2

E1

β2 = 0, 5β1 (T2 = 2T1 )

p4

E4

p3

p4

E3

p2

E2

p1

E1

p3

p2

p1

Fig. 2.4: The considered system has four microstates of energies E1 to E4 . The abscissae are proportional to the probabilities of occurrence of one of the microstates, for two diﬀerent values of β (or of the temperature). If the total number of microstates of system A corresponding to the same energy E is W (E), then the probability that system A, coupled to A , has the energy E, is equal to : p(E) = C W (E) exp(−β0 E)

(2.33)

where W (E) is the degeneracy (the number of ways to realize this energy) of the state of energy E. The probability p(E) is the product of the function W (E), very rapidly increasing with energy for a macroscopic system, by a ˜ decreasing exponential : the probability has a very steep maximum for E = E.

2.4.2

Energy, with Fixed Average Value

We now show that another condition, i.e., the constraint of a given average value E of the energy of a system A, leads to a probability law analogous to

Canonical Ensemble

43

the one just found. The only diﬀerence is that the value of β, now a parameter, should be adjusted so that the average of energy indeed coincides with the given value E. We have to maximize the statistical entropy of A while the value E is given. We thus have to realize SA = −kB

pi ln pi maximum, under the constraints

i

⎧ ⎪ pi = 1 ⎪ ⎨ i ⎪ pi Ei = E ⎪ ⎩

(2.34)

i

A general mathematical method allows one to solve this type of problem, it is the method of Lagrange multipliers explained in Appendix 2.1 : parameters are introduced, which are the Lagrange multipliers (here β). These parameters are adjusted at the end of the calculation, in order that the average values be equal to those given by the physical properties of the system, here the value of the average energy. One thus obtains for the probability of occurrence of the microstate of energy Ei : exp(−βEi ) pi = exp(−βEi )

(2.35)

i

that is, an expression of the “Boltzmann-factor” type (2.32) but here β does not characterize the temperature of a real heat reservoir. It is rather determined by the condition that the average energy of the system A should be the one given by the problem conditions, that is :

Ei exp(−βEi )

i

exp(−βEi )

= E

(2.36)

i

This method of the Lagrange multipliers is followed in some Statistical Physics courses to obtain the general expression of the Boltzmann factor (for example, those given at Ecole polytechnique by R. Balian, E. Brézin, or A. Georges and M. Mézard).

44

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

Partition Function Z

2.4.3

The probability for the system in contact with a heat reservoir to be in a state of energy Ei is pi =

e−βEi 1 , withβ = ZN kB T

The term ZN =

(2.37)

e−βEi

(2.38)

i

which allows one to norm the Boltzmann factors is called the “canonical partition function for N particles” (in German “Zustandssumme,” which means “sum over the states”). Formulae (2.37) and (2.38) are among the most important ones of this course ! The complete statistical information on the considered problem is contained in its partition function, since from Z and its partial derivatives all the average values of the physical macroscopic parameters of the system can be calculated. We will show this property now on the examples of the average energy and the energy ﬂuctuation around its average (for the entropy see § 2.4.5). Indeed, the average energy is given by E =

pi Ei =

i

Ei

i

exp(−βEi ) Z

(2.39)

∂ ln Z 1 ∂Z =− =− Z ∂β ∂β It appears here that ln Z is extensive like E.

To determine the energy ﬂuctuation around its average value, one calculates (∆E)2 = E − E2 = E 2 − E2

(2.40)

A procedure similar to that of (2.39) is followed to ﬁnd E 2 : E 2 =

pi Ei2 =

i

(∆E)2 =

1 ∂2Z 1 − 2 Z ∂β 2 Z

Ei2

i

Å

∂Z ∂β

exp(−βEi ) 1 ∂2Z = Z Z ∂β 2 ã2 =

∂ ∂β

Å

1 ∂Z Z ∂β

ã =

(2.41) ∂ 2 ln Z ∂β 2

(2.42)

One veriﬁes here that, since ln Z varies as the number of particles of the system and β is intensive, ∆E varies as N 1/2 and ∆E/E as N −1/2 (in agreement

Canonical Ensemble

45

with the result of § 2.3.2), which corresponds to an extremely small value in a macroscopic system : for N = 1023 , ∆E/E is of the order of 10−11 ! Note 1 : A demonstration (outside the framework of the present course, see, for example, R. Balian chapter 5 § 5) can be done of the equivalence between all the statistical ensembles in the case of macroscopic systems : it is valid in the “thermodynamical limit,” which consists in simultaneously having several parameters tending toward inﬁnity : the volume Ω (which means in particular that it is very much larger than typical atomic volumes), the particle number N and the other extensive parameters, while keeping constant the particle density N/Ω and the other intensive parameters. In such a limit, it is equivalent to impose an exact value of a physical parameter (like the energy in the microcanonical ensemble) or an average value of the same parameter (the energy in the canonical ensemble). In fact, the technique is simpler in the canonical ensemble, where ln Z is calculated on all the states without any restriction, than in the microcanonical ensemble, where the calculation of ln W (E) requires the limitation to the range of energies between E and E + δE. Besides, for macroscopic systems one understands that the predicted physical results, using either the microcanonical or the canonical statistical ensemble, cannot be distinguished through measurements. Note 2 : The value of ZN “gives a hint” at the number of microstates that can be achieved at the experiment temperature. Indeed at very low temperature, where βEi → ∞, only the fundamental state E0 is realized and ZN = 1. On the other hand, at high temperature where βEi → 0, many terms are of the order of unity, corresponding to practically equally likely states.

2.4.4

Entropy in the Canonical Ensemble

We just showed that, using the partition function, we can calculate the average energy of system A at temperature T , together the ﬂuctuation around this average energy. We now have to express the statistical entropy of A in these conditions. Here the probabilities of occurrence of the diﬀerent microstates are not equal, since they depend of their respective energies through Boltzmann factors. Thus the deﬁnition of the statistical entropy to be used is no longer that of Boltzmann, but rather the one of Gibbs : SA = −kB

i

pi ln pi

(2.43)

46

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

with pi =

1 −βEi e , pi = 1 Z i

(2.44)

Consequently, SA = −kB

pi (− ln Z − βEi )

(2.45)

i

SA = kB (ln Z + βE) = kB ln Z + kB βU

(2.46)

The average energy E of the macroscopic system is identiﬁed to the internal energy U of Thermodynamics (see § 3.2). Note that, for a large system in which the energy ﬂuctuation is relatively very small around E, the probability of occurrence of this energy value is very close to unity. The number of occurrences of this average energy is W (E), so that the Gibbs entropy (2.43) practically reduces to the Boltzmann entropy (2.10) for this value E.

2.4.5

Partition Function of a Set of Two Independent Systems with the Same β Parameter (Same T )

In § 2.1 we saw that, in the N -particle problems solved in this book, we always decompose the hamiltonian of the total system into a sum of hamiltonians for individual particles ; a particle may have several degrees of freedom, whence another sum of hamiltonians. The energy of a microstate is thus a sum of energies and now we see the consequence of this property on the partition function of such a system in thermal equilibrium. This situation is schematized by simply considering a system A made up of two independent subsystems A1 and A2 , both in contact with a heat reservoir at temperature T . A particular microstate of A1 has the energy E1i , while the energy of A2 is E2j and the energy of A is Eij . Then ⎧ Eij = E1i + E2j ⎪ ⎪ ⎪ ⎪ ⎪ ⎪ 1 ⎪ ⎪ Z = Z1 Z2 = exp{−β(E1i + E2j )} , with β = ⎪ ⎪ kB T ⎨ ij ln Z = ln Z1 + ln Z2 ⎪ ⎪ ⎪ ⎪ ⎪ ⎪ E = E1 + E2 ⎪ ⎪ ⎪ ⎪ ⎩ S = S1 + S2

(2.47)

Canonical Ensemble

47

This property is very often used : when the energies of two independent systems at the same temperature sum, the corresponding partition functions multiply. Consequently, in the case of distinguishable independent particles with several degrees of freedom, one will separately calculate the partition functions for the various degrees of freedom of an individual particle ; then one will multiply the factors corresponding to the diﬀerent degrees of freedom and particles to obtain the partition function Z of the total system. Finally, in the special case of the ideal gas, one will introduce the factor CN into Z (see § 4.4.3 and 6.7). Take the example of the vibrations of atoms around their equilibrium positions, owing to thermal motion, in a solid in thermal equilibrium at temperature T : this is a very classical exercise and here we only sketch its solution. In the model proposed by Einstein (1907), one assumes that the atoms, in number N , are points and that they are all attracted toward their respective equilibrium position with the same restoring constant, that is, with the same frequency ω. The value of ω depends of the mechanical properties of the solid, in particular of its stiﬀness. Then the hamiltonian for a particular atom in motion is given by 1 ˆ 2i ˆ xi + ˆhyi + h ˆ zi ˆi = p + mω 2ri2 = h h 2m 2

(2.48)

It is the sum of three hamiltonians of the “harmonic-oscillator” type, identical ˆ xi relative to the for each coordinate. The eigenvalues of the hamiltonian h coordinate x of site i are

1 n = nxi + ω Exi 2

(2.49)

The contribution to the partition function of this degree of freedom is equal to exp(−β(nxi + 1/2)ω) (2.50) zxi = n

The partition function for the total system is the product of 3N terms similar to this one. The average energy is deduced using (2.39), it is equal to 3N times the average value for the coordinate x of site i. The parameter accessible to experiment is the speciﬁc heat at constant volume Cv , deﬁned as the derivative of the average energy with respect to temperature. The above solution gives for this lattice contribution Cv =

3N kB (Θ/T )2 ω , taking =Θ Θ/T 2 kB (e − 1)

This expression is sketched on Fig. 2.5.

(2.51)

48

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

Cv /3N kB 1

O

1

T /Θ

Fig. 2.5: Variation of Cv versus temperature, according to the Einstein model. The dots are experimental data for diamond, the solid line is calculated for Θ = 1320 K (after A. Einstein, Annalen der Physik vol.22, p.186 (1907)). Its high temperature limit Cv = 3N kB = 25 J.K−1mol−1 is independent of ω, thus universal : this is the Dulong and Petit law (1819), valid for metals at room temperature. The low temperature limit Cv → 0 for T → 0 is in agreement with experiment. However, the above law of variation of Cv is not experimentally veriﬁed. The Einstein model is improved by introducing a repartition of characteristic frequencies in a solid instead of a single ω : this is the Debye model (1913), in which Cv is proportional to T 3 at very low temperatures. The agreement with experiment then becomes excellent for insulators, whereas for metals an additional term is due to the mobile electron contribution to the speciﬁc heat (see § (7.2.2)).

2.5

Exchange of Both Energy and Particles : “Grand Canonical Ensemble”

Here we consider two coupled systems A and A , separated by a porous partition through which energy and particles can be exchanged, the combined system A0 being isolated from its surroundings (Fig. 2.6). An example of such a situation is water molecules under two phases, liquid and gas, contained in

Grand Canonical Ensemble

∂S = −kB α ∂N A ∂S = kB β ∂E

49

E

∂S = −kB α ∂N A

N

∂S = kB β ∂E

Fig. 2.6 : The two systems A and A exchange energy and particles. a thermally insulated can of ﬁxed volume, the water molecules belonging to either phase. One can also study hydrogen molecules as a free gas in equilibrium with hydrogen molecules adsorbed on a solid palladium catalyst surface together with the total system being thermally insulated and the total number of molecules ﬁxed. We have to state that the entropy is maximum at equilibrium, following an approach very similar to the one of 2.3.1.

2.5.1

Equilibrium Condition : Equality of Both Temperatures and Chemical Potentials

One writes that the maximum of entropy for the combined system A0 = A+A takes place at constant total energy and ﬁxed total number of particles, i.e., SA0 (E0 , N0 ) = SA (E, N ) + SA (E , N ) maximum under the two constraints

(2.52)

⎧ ⎨ E0 = E + E

(2.53)

⎩ N = N + N 0 Let us express the condition of entropy maximum of A0 : SA0 (E0 , N0 ) = SA (E, N ) + SA (E0 − E, N0 − N )

(2.54)

With respect to an energy exchange between the two systems : Å ã ∂SA0 (E, N ) =0, ∂E N Å i.e.,

∂SA0 (E, N ) ∂E

ã

Å =

N

∂SA (E, N ) ∂E

ã

Å +

N

∂SA (E0 − E, N0 − N ) ∂E

ã =0 N

(2.55)

50

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

With respect to a particles exchange : Å ã ∂SA0 (E, N ) =0 ∂N E Å

∂SA0 (E, N ) i.e., ∂N

ã

Å =

E

∂SA (E, N ) ∂N

ã

Å + E

∂SA (E0 − E, N0 − N ) ∂N

ã = 0. E

(2.56) The thermal equilibrium condition (2.55) again provides the results of § 2.3.1, that is, β = β . The condition (2.56) on N implies that ã ã ã Å Å Å ∂SA (E0 − E, N0 − N ) ∂SA (E , N ) ∂SA (E, N ) =− = ∂N ∂N ∂N E E E (2.57) The second equilibrium condition is the equality, in the systems A and A in contact, of the partial derivatives of the entropy versus the number of particles, all the other parameters being kept constant. This derivative will be written Å ã ∂S = −kB α (2.58) ∂N Ω,E where kB is the Boltzmann constant. It will be seen in § 3.5.1 that the parameter α is related to the chemical potential µ through α = βµ = µ/kB T . The equilibrium condition between systems A and A thus simultaneously implies ⎧ ⎪ ⎨ β = β , i.e., T = T (2.59) ⎪ ⎩ α = α , i.e., µ = µ T T that is, the equality of both the temperatures and the chemical potentials in both systems. Thus in equilibrium the chemical potential of hydrogen is the same, whether in gas phase or adsorbed on the catalyst, and the temperatures are the same in both phases.

2.5.2

Heat Reservoir and Particles Reservoir ; Grand Canonical Probability and Partition Function

Similarly to the approach of § 2.4, one now considers that the system A under study is macroscopic, but much smaller that the system A (Fig. 2.7). Consequently, the parameters α and β of system A are almost the same as those

Grand Canonical Ensemble

51

of the combined system A0 , and are very little modiﬁed when system A varies in energy or in number of particles : with respect to A the system A behaves like a heat reservoir and a particle reservoir, it dictates both its temperature β = 1/kB T and its chemical potential α = µ/kB T to A.

α

E A

A N

β

Fig. 2.7: The system A is in contact with the heat reservoir and particles reservoir A , which dictates its parameters α and β to A. This situation in Statistical Physics, in which the system under study A is coupled to a heat reservoir that dictates its value of β, i.e., its temperature, and to a particle reservoir that gives the value of α, is called the “grand canonical ensemble.”

2.5.3

Grand Canonical Probability Law and Partition Function

Now both the energy Ei and the number of particles Ni of the considered system enter into the Boltzmann-type probability pi . Indeed, consider the probability pi to reach a particular microstate i of A, of energy Ei and particle number Ni . It is related to the ability for the reservoir to realize this physical situation at given energy and number of particles, that is, to the number Wres (E0 − Ei , N0 − Ni ) of equally likely microstates, and thus to the reservoir entropy, through the Boltzmann entropy relation. Now Å ã ∂Sres Sres (E0 − Ei , N0 − Ni ) = SA0 (E0 , N0 ) − Ei ∂E E0 ,N0 Å ã (2.60) ∂Sres − Ni + ... ∂E E0 ,N0 thus ln Wres (E0 − Ei , N0 − Ni ) = ln Wres (E0 , N0 ) − Ei β0 + Ni α0 + . . .

(2.61)

The probability of occurrence of this microstate is proportional to Wres , i.e., p(Ei , Ni ) =

1 exp(−β0 Ei + α0 Ni ) ZG

(2.62)

52

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

which deﬁnes the grand partition function ZG , the quantity which norms the probabilities. It is now a sum over all energies and numbers of particles that can be achieved ; this sum can also be regrouped according to the diﬀerent numbers of particles : e−βEi +αNi = eαN ZN (2.63) ZG = i

N

The canonical partition function ZN for N particles, as deﬁned in (2.38), e−βEi (2.64) ZN = i

thus appears in ZG with a weight equal to exp αN . (Note that in (2.63) the sum is performed over all microstates i with an arbitrary number of particles, whereas in (2.64) the sum is only over the states with N particles, with N ﬁxed.)

2.5.4

Average Values

Using p(Ei , Ni ) one calculates the average energy and the average number of particles of system A : this is conveniently performed from the partial derivatives of ZG . The approach is the same as in (2.39) : E =

Ei p(Ei , Ni ) =

i

Ei

i

∂ ln ZG exp(−βEi + αNi ) 1 ∂ZG =− =− ZG ZG ∂β ∂β (2.65)

N =

Ni p(Ei , Ni ) =

i

i

Ni

∂ ln ZG exp(−βEi + αNi ) 1 ∂ZG = =+ ZG ZG ∂α ∂α (2.66)

To determine the energy ﬂuctuations around its average value, one proceeds in a way similar to the calculation in the canonical ensemble (eqs. (2.41) and (2.42)). Let us calculate the ﬂuctuation of the particle number in system A : N 2 =

i

Ni2 p(Ei , Ni ) =

1 ∂ 2 ZG ZG ∂α2

1 ∂ 2 ZG − (∆N )2 = N 2 − N 2 = ZG ∂α2 Å ã 1 ∂ZG ∂ 2 ln ZG ∂ = = ∂α ZG ∂α ∂α2

Å

(2.67) 1 ∂ZG ZG ∂α

ã2 (2.68)

Other Statistical Ensembles

53

Since ln ZG is extensive as it increases like E or N (see (2.65) and (2.66)), (∆N )2 varies proportionaly with the size of the system, or with its average number of particles. Consequently, ∆N/N is of the order N −1/2 , it is thus very small for a macroscopic system. Thus by diﬀerentiating ZG one obtains the average physical parameters E and N and their ﬂuctuations around these averages : in fact ZG , which expresses the probabilities, contains the complete information on the system.

2.5.5

Grand Canonical Entropy

Using the Gibbs deﬁnition S = −kB

pi ln pi

(2.69)

i

one calculates the entropy in the grand canonical ensemble. Here pi =

1 −βEi +αNi e , pi = 1 ZG i

(2.70)

Consequently, S = −kB

pi (− ln ZG − βEi + αNi )

(2.71)

i

S = kB (ln ZG + βE − αN ) = kB ln ZG + kB βU − kB αN

(2.72)

In the present deﬁnition of the grand canonical ensemble, we have considered a macroscopic system in contact with a heat reservoir and a particle reservoir. The grand canonical ensemble is also used when the system under study is speciﬁed by the average values of both its energy and its particle number and one then needs to determine β and α = βµ : the argument is the extension of the one in § 2.4.2. We will use the grand canonical ensemble in particular for the study of the Quantum Statistics : the Pauli principle introduces limitations to the occupation of the quantum states, which will be more conveniently expressed in this ensemble.

2.6

Other Statistical Ensembles

As you imagine, one can generalize to various physical situations the approach consisting of considering two weakly coupled systems A and A , the combined system being isolated, i.e., the total parameters being ﬁxed. In particular one

54

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

can study systems A and A exchanging both energy and volume, the total volume Ω0 = Ω + Ω being ﬁxed. The equilibrium condition now implies that the partial derivatives of the entropy ∂S ∂E

N,Ω

= kB β

and

∂S ∂Ω

N,E

= kB γ

(2.73)

take the same value in both systems. Thermodynamics will show, in § 3.5.1, that kB γ = P/T where P is the pressure, so that the new equilibrium condition will be ® T = T (2.74) P = P that is, in equilibrium the systems A and A will have the same temperature and the same pressure. If A is much larger than A, it will give its pressure to A. The chemistry experiments performed in isothermal and isobar conditions (for example, at room temperature and under the atmospheric pressure) are in such a condition. Then Å ã P Ωi 1 1 Ei − exp(−βEi − γΩi ) = exp − (2.75) p(Ei , Ωi ) = ZT,P ZT,P kB T kB T i represents the microstate in which the system has the energy Ei and occupies the volume Ωi . A new partition function ZT,P is deﬁned by exp(−βEi − γΩi ) (2.76) ZT,P = i

(In fact the sum over the volume, a continuous variable, is rather an integral.) The entropy is deduced : S = −kB p(Ei , Ωi )(− ln ZT,P − βEi − γΩi ) (2.77) i

S = kB (ln ZT,P + βE + γΩ)

(2.78)

Summary of Chapter 2 In this chapter the general method for solving problems in Statistical Physics is presented. – First one needs to determine the quantum states of the system. In the cases we will consider, the system hamiltonian can be decomposed into a sum of similar one-particle terms ˆi ˆ = H h i

Until chapter 4, the particles will be assumed to be distinguishable. – Only in a second step does one choose a statistical ensemble adapted to the physical conditions of the problem (or the most convenient to solve the problem, since for macroscopic systems one cannot measure the ﬂuctuations resulting from the choice of a speciﬁc ensemble rather than another one). One expresses that the entropy is maximum in equilibrium, under the constraints set by the physical conditions. According to the chosen ensemble, a mathematical quantity deduced from the probabilities allows one to obtain the macroscopic physical parameters that can be measured in experiments : – the microcanonical ensemble is used to treat the case of an isolated system : the number of microstates W (E) all producing the same macrostate of energy in the range between E and E + δE, allows the entropy to be obtained S = kB ln W (E) – the canonical ensemble is adapted to the case of a system with a ﬁxed number of particles N , in contact with a heat reservoir at temperature T . The probability to be in a microstate of energy Ei is then given by the Boltzmann factor : e−βEi 1 , with β = pi = ZN kB T 55

56

Summary of Chapter 2

The canonical partition function for N particles is deﬁned by ZN = e−βEi i

It is related to the average energy through E = −

∂ ln Z 1 ∂Z =− Z ∂β ∂β

– the grand canonical ensemble corresponds to the situation of a system in contact with a heat reservoir and a particle reservoir. The grand canonical partition function is introduced : ZG = eαN ZN N

The probability to be in a microstate of both energy Ei and number of particles Ni is given by p(Ei , Ni ) =

1 exp(−βEi + αNi ) ZG

The average energy and the average number of particles are then deduced : E = −

∂ ln ZG 1 ∂ZG ∂ ln ZG 1 ∂ZG =− , N = = ZG ∂β ∂β ZG ∂α ∂α

Appendix 2.1 Lagrange Multipliers One looks for the extremum of a function with several variables, satisfying constraints. Here we just treat the example of the entropy maximum with a ﬁxed average value of the energy, i.e., pi ln pi maximum, under the constraints S = −kB i

⎧ pi = 1 ⎪ ⎨ i ⎪ pi Ei = E ⎩

(2.79)

i

The process consists in introducing constants, the Lagrange multipliers, here λ and β, the values of which will be determined at the end of the calculation. Therefore one is looking for the maximum of the auxiliary function : F (p1 , . . . , pi , . . . ) = − pi ln pi + λ( pi − 1) − β( Ei pi − E) (2.80) i

i

i

which expresses the above conditions, with respect to the variables pi . This is written ∂F = −(ln pi + 1) + λ − βEi = 0 ∂pi

(2.81)

ln pi = −βEi + λ − 1

(2.82)

pi = (eλ−1 ) e−βEi

(2.83)

The solution is

Consequently,

57

58

Appendix 2.1

The constants λ and β are then obtained by expressing the constraints : • pi = 1 , whence i

eλ−1

e−βEi = 1

i

eλ−1 = •

1 e

−βEi

=

(2.84) 1 Z

(2.85)

i

pi Ei = E , i.e.,

(2.86)

i

1 Ei e−βEi = E Z i

(2.87)

which is indeed expression (2.36). Note that, if the only constraint is Σpi = 1, the choice of the pi maximizing S i

indeed consists in taking all the pi equal ; this is what is done for an isolated system using the Boltzmann formula.

Chapter 3

Thermodynamics and Statistical Physics in Equilibrium In Chapter 1 we stated the basic principles of Statistical Physics, which rely on a detailed microscopic description and on the hypothesis, for an isolated system, of the equiprobability of the diﬀerent microscopic occurrences of a given macroscopic state ; the concept of statistical entropy was also introduced. In Chapter 2 the study was extended to other systems in equilibrium, in which the statistical entropy is maximized, consistently with the constraints deﬁned by the physical conditions of the problem. It is now necessary to link this statistical entropy S and the entropy which appears in the second law of Thermodynamics and that you became familiar with in your previous studies. It will be shown in this chapter that these two entropies are identical. Thus the purpose of this chapter will ﬁrst be (§ 3.1 to 3.4), mostly in the equilibrium situation, to rediscover the laws of Thermodynamics, which rule the behavior of macroscopic physical properties. This will be done in the framework of the hypotheses and results of Statistical Physics, as obtained in the ﬁrst two chapters of this book. Then we will be able to use at best the diﬀerent thermodynamical potentials (§ 3.5), each of them being directly related to a partition function, that is, to a microscopic description. They are convenient tools to solve the Statistical Physics problems. We are going to state the successive laws of Thermodynamics and to see how each one is interpreted in the framework of Statistical Physics. The example chosen in Thermodynamics will be the ideal gas, the properties of which you 59

60

Chapter 3. Thermodynamics and Statistical Physics

studied in previous courses. Its study in Statistical Physics will be presented in the next chapter 4.

3.1

Zeroth Law of Thermodynamics

According to this law, if two bodies at the same temperature T are brought into thermal contact, i.e. can exchange energy, they remain in equilibrium. In Statistical Physics, we saw in (2.3.1) that if two systems are free to exchange energy, the combined system being isolated, the equilibrium condition is that the partial derivative of the statistical entropy versus energy should take the same value in each system : ∂S ∂S = kB β, = kB β ; β = β (3.1) ∂E ∂E It was suggested that this partial derivative of the entropy is related to its temperature, as will be more precisely shown below in § 3.3.

3.2

First Law of Thermodynamics

Its statement is as follows : Consider an inﬁnitesimal process of a system with a ﬁxed number of particles, exchanging an inﬁnitesimal work δW and an inﬁnitesimal amount of heat δQ with its surroundings. Although both δW and δQ depend on the sequence of states followed by the system, the sum dU = δW + δQ

(3.2)

only depends on the initial state and the ﬁnal state, i.e., it is independent of the process followed. This states that dU is the diﬀerential of a state function U , which is the internal energy of the system. A property related to this ﬁrst law is that the internal energy of an isolated system remains constant, as no exchange of work or heat is possible : dU = 0

(3.3)

In this Statistical Physics course, we introduced the energy of a macroscopic system : in the microcanonical ensemble it is given exactly ; in the canonical or grand canonical ensemble its value is given by the average pi Ei (3.4) E = U = i

First Law of Thermodynamics

61

where pi is the probability of occurrence of the state of energy Ei . The diﬀerential of the internal energy pi dEi + Ei dpi (3.5) dU = i

i

consists in two terms that we are going to identify with the two contributions in dU , i.e., δW and δQ. For this purpose, two particular processes will ﬁrst be analyzed, in which either work (§ 3.2.1) or heat (§ 3.2.2) is exchanged with the surroundings. Then the case of a quasi-static general process will be considered (§ 3.2.3).

3.2.1

Work E

E

E4 + dE4 E4 E3 + dE3 E3

E2 E2 + dE2

E1 + dE1 E1 xj

xj + dxj

Fig. 3.1: Modiﬁcation of the energies of the considered system when the external parameter varies from xj to xj + dxj . The quantum states of a system are the solutions of a hamiltonian which depends of external parameters xj characteristic of the system : for example, you learnt in Quantum Mechanics, and will see again in chapter 6, that a gas molecule, free and conﬁned within a ﬁxed volume Ω, assumes for energy

62

Chapter 3. Thermodynamics and Statistical Physics

eigenvalues 2 k 2 /2m, where the allowed values of k depend of the container volume like Ω−1/3 ; as for the probabilities of realizing these eigenstates, they are determined by the postulates of Statistical Physics. Mechanical work will inﬂuence these external parameters : for example, the motion of a piston modiﬁes the volume of the system, thus shifting its quantized states. More generally, if only the external parameters of the hamiltonian are modiﬁed from xj to xj + dxj , the energy of the eigenstates is changed and a particular eigenvalue Ei of the hamiltonian becomes Ei + dEi in such a process (Fig. 3.1). The variation of the average energy of the system is then ∂U dxj = −Xj dxj ∂xj

(3.6)

where Xj is the generalized force corresponding to the external parameter xj . In the special case of a gas, the variation of internal energy in a dilatation dL is equal to ∂U dL = −F dL = −P SdL = −P dΩ ∂L

(3.7)

Here P is the external pressure that acts on the system surface S ; it is equal to the pressure of the system in the case of a reversible process (Fig. 3.2). This internal energy variation is identiﬁed with the inﬁnitesimal work received by the gas. In such a situation the variation dL of the characteristic length modiﬁes the accessible energies. P dΩ

S

dL

Ω

Fig. 3.2 : Dilatation of the volume Ω submitted to the external pressure P . In general the variation of average energy resulting from the variation of the external parameters is identiﬁed with the inﬁnitesimal work δW = − Xj dxj (3.8) j

and corresponds to the term

i

pi dEi

First Law of Thermodynamics

3.2.2

63

Heat

Assume that energy is brought into the system, while no work is delivered : the external parameters of the hamiltonian are unchanged, the energy levels Ei remain identical. The average energy of the system increases, now due to higher probabilities of occurrence of the system states of higher energy. The probabilities pi are modiﬁed and now the variation of internal energy is identiﬁed with the inﬁnitesimal heat absorbed : δQ = Ei dpi (3.9) i

A way to conﬁrm this identiﬁcation of

Ei dpi is to consider a very slow pro-

i

cess in which no heat is exchanged with the surroundings, that is, the system is thermally insulated : this is an adiabatic ∗ process. In Thermodynamics, such a process is done at constant entropy. In Statistical Physics, it can also be shown that the statistical entropy is conserved in this process : pi ln pi = constant (3.10) S = −kB i

Consequently,

Now as

dpi ln pi +

i

pi = 1, one has

i

pi

i

dpi =0 pi

(3.11)

dpi = 0, whence

i

ln pi dpi = 0

(3.12)

i

One replaces ln pi by its expression in the canonical ensemble ln pi = −βEi − ln Z and deduces

Ei dpi = 0

(3.13)

(3.14)

i

which is indeed consistent with the fact that the heat δQ exchanged with the surroundings is zero in an adiabatic process.

3.2.3

Quasi-Static General Process

Consider an inﬁnitesimal quasi-static ∗ process of a system, that is, slow enough to consist in a continuous and reversible ∗ sequence of equilibrium states.

64

Chapter 3. Thermodynamics and Statistical Physics

In the diﬀerential dU one still identiﬁes the work with the term arising from the variation of the hamiltonian external parameters in this process. The complementary contribution to dU is identiﬁed with δQ, the heat exchanged with the surroundings in the process.

3.3

Second Law of Thermodynamics

Let us recall its statement : Let us consider a system with a ﬁxed number N of particles, in thermal contact with a heat source at temperature T . The system exchanges an amount of heat δQ with this source in an inﬁnitesimal quasi-static process. Then the entropy variation dSthermo of the considered system is equal to dSthermo =

δQ T

(3.15)

Although δQ depends on the particular path followed, the extensive quantity dSthermo is the diﬀerential of a state function, which thus only depends on the initial and ﬁnal states of the considered process. Besides, in a spontaneous, i.e. without any contribution from the surroundings, and irreversible process, the entropy variation dSthermo is positive or zero. The general deﬁnition of the entropy in Statistical Physics is that of Gibbs : S = −kB

pi ln pi

(3.16)

i

Let us ﬁrst examine the interpretation of an irreversible process in Statistical Physics. An example is the free expansion of a gas into vacuum, the so-called “Joule-Gay-Lussac expansion” : a N -molecule gas was occupying a volume Ω (Fig. 3.3a), the partition is removed, then the gas occupies the previously empty volume Ω (Fig. 3.3b). The combined system is isolated and removing the partition requires a negligible amount of work. The accessible volume becomes Ω+Ω and it is extremely unlikely that the molecules again gather in the sole volume Ω. In such a process, which is irreversible in Thermodynamics, the accessible volume for the particles has increased, that is, the constraint on their positions has been removed ; the number of cells of the one-particle phase space (r, p) has increased with the accessible volume in this process. Let us now consider an inﬁnitesimal reversible process, corresponding to a mere modiﬁcation of the probabilities of occurrence of the same accessible

Second Law of Thermodynamics

65

Ω

Ω

Ω

Ω

a)

b)

Fig. 3.3 : Joule-Gay-Lussac expansion into vacuum. microstates. dS is obtained by diﬀerentiating (3.16) : pi ln pi dpi + dpi dS = −kB pi i i ln pi dpi = −kB i

since, as probabilities are normalized,

(3.17)

dpi = 0.

i

To be more precise, let us work in the canonical ensemble, in which the system is submitted to this process at constant number of particles, remaining at every time t in thermal equilibrium with a heat reservoir at the temperature T (t). The probability to realize the microstate of energy Ei is equal to (see (2.37)) : pi (t) =

e−β(t)Ei (t) , i.e., ln pi (t) = −β(t)Ei (t) − ln Z(t) Z

Then dS = kB β(t)

Ei (t)dpi

(3.18)

(3.19)

i

From the above interpretation of the inﬁnitesimal heat δQ, (3.19) can also be written dS = kB β(t)δQ

(3.20)

This expression is to be compared to (3.15). The statistical entropy deﬁned by (3.16) and the thermodynamical entropy (3.15) can be identiﬁed under the condition that kB β, deﬁned by (3.1) as the partial derivative of the statistical 1 entropy S versus energy, is identical to , the reciprocal of the thermodynaT mical temperature : kB β ≡

1 T

(3.21)

66

Chapter 3. Thermodynamics and Statistical Physics

Indeed the integrand factor, which allows one to transform δQ into the total diﬀerential dS, is unique to a constant, which justiﬁes this identiﬁcation.

3.4

Third Law of Thermodynamics

The second law allows one to calculate entropy variations. From the third law, or Walther Nernst (1864-1941)’s law, the absolute value of S can be obtained : When the absolute temperature T of a system tends to zero, its entropy tends to zero. From a microscopic point of view, when the temperature vanishes, the particles gather into the accessible states with the lowest energies. If the minimum energy corresponds to a single state E0 , the probability of occupation of this fundamental state tends to unity while those for the other states vanish. Then the microscopic description becomes certain and the entropy vanishes : for example, when the temperature tends to zero in a monoatomic ideal gas, one expects the molecules to be in a state of kinetic energy as small as possible (yet diﬀerent from zero, because of the Heisenberg uncertainty principle) ; in the case of electrons in a solid, as will be studied in chapter 7, the fundamental state is nondegenerate too. Let us now consider a material which has two equivalent states at low temperature, with the same minimal energy, for example, two allotropic crystalline species, each occurring with the probability 1/2. When T tends to zero, the corresponding entropy tends to S0 = kB ln 2

(3.22)

This quantity is diﬀerent from zero, yet it is not extensive so that S0 /N becomes negligible. This is very diﬀerent from the situation at ﬁnite T where S/N is ﬁnite and S extensive. Thus the statement of the Nernst law consistent with Statistical Physics is that S/N tends to zero for N large when the temperature tends to the absolute zero. Notice that, in the case of the paramagnetic solid at zero external magnetic ﬁeld and very low temperature, one would expect to ﬁnd S/N = kB ln 2, in contradiction with the Nernst law. In fact, this spin system cannot be strictly in zero magnetic ﬁeld because of the nuclear magnetic interactions that are always present (but are not very eﬃcient at higher temperature where other phenomena imply much larger amounts of energy) : under the eﬀect of these

The Thermodynamical Potentials ; the Legendre Transformation

67

nuclear interactions the system is no longer at zero magnetic ﬁeld and becomes ordered when T tends to zero. The present sections have put in correspondence the laws of Thermodynamics and the statistical approach of chapters 1 and 2. Their conclusion is that the description based on Statistical Physics and its postulates is indeed equivalent to the three laws of Thermodynamics.

3.5

The Thermodynamical Potentials and their Diﬀerentials ; the Legendre Transformation

Now that we have identiﬁed the statistical entropy to the thermodynamical entropy, we can take advantage of our background in Thermodynamics to introduce the thermodynamical potential adapted to the considered physical situation, and relate it to the statistical description and to the physical parameters. The thermodynamical potential will be a function of the parameters determined by the experimental conditions and in equilibrium it will be extremum.

3.5.1

Isolated System

Here the entropy S is the thermodynamical potential to consider. During a spontaneous evolution toward equilibrium, S increases and reaches its maximum in equilibrium. In the microcanonical description, the entropy is related to the number W of microstates of energies lying in the range between E and E + dE by S = kB ln W

(3.23)

dS and dU are linked through the diﬀerential expressions of the ﬁrst and second laws of Thermodynamics : dS =

dU δW δQ = − T T T

(3.24)

so that 1 = T

Å

∂S ∂U

ã (3.25) N,Ω

68

Chapter 3. Thermodynamics and Statistical Physics

In the case of the ideal gas where δW = −P dΩ

dS =

(3.26)

dU P dΩ + T T

(3.27)

The diﬀerential expression (3.27) shows that S, calculated for a ﬁxed number N of particles, is a function of the variables U and Ω, both extensive. The pressure is deduced from the partial derivative of S versus the volume : Å ã ã Å P ∂S kB ∂ ln W (U ) = = (3.28) T ∂Ω N,U ∂Ω N,U Note : we suggested in the Appendix 3 of chapter 1 that, in the case of free particles, the number of microstates associated with the energy E was of the form W (E) ∝ ΩN χ(E)

(3.29)

so that P = T

Å

kB ∂ ln W (U ) ∂Ω

ã = N,U

kB N Ω

P Ω = N kB T

(3.30) (3.31)

Here we ﬁnd the ideal gas law, that will be again established in the next chapter. After having studied a given system, we now consider another similar system, with the same energy U and under the same volume Ω, which only diﬀers by its number of particles equal to N + dN . Its entropy is now S(U, Ω, N ) + dS, and we write µ dS = − dN T

(3.32)

This deﬁnes the chemical potential µ, related to the entropy variation of the system when an extra particle is added, at constant internal energy and volume. The chemical potential is an intensive parameter. For a classical ideal gas, if N increases, one expects S to increase, so that µ is negative. Then, including the dependence in the particles number, the diﬀerential of S for the ideal gas is written dS =

dU P dΩ µdN + − T T T

(3.33)

The Thermodynamical Potentials ; the Legendre Transformation

69

S is a function of extensive variables only ; the partial derivative of S versus N is equal to ã Å µ ∂S (3.34) =− ∂N Ω,U T We previously showed in § 2.5.1 that if the two systems are exchanging particles, in equilibrium ∂S = −kB α ∂N takes the same value in both systems.

(3.35)

These two partial derivatives of the entropy versus the number of particles can be identiﬁed, so that µ (3.36) α= kB T Besides, from the diﬀerential (3.33) of S expressed for the ideal gas one deduces the expression of the diﬀerential of its energy U : dU = −P dΩ + T dS + µdN

(3.37)

The last term can be interpreted as a “chemical work.”

3.5.2

System with a Fixed Number of Particles, in Contact with a Heat Reservoir at Temperature T

We are going to show that the corresponding thermodynamical potential is the Helmholtz free energy F , which is minimum in equilibrium and is related to the N -particle canonical partition function ZN . Indeed the general deﬁnition (3.16) of S applies, with now probabilities pi equal to pi =

1 exp(−βEi ) ZN

Then i

pi ln pi =

i

=−

Å pi

and β =

Ei − kB T

1 kB T

(3.38)

ã − ln ZN

U − ln ZN kB T

(3.39)

Consequently, in the canonical ensemble, U + kB ln ZN T U − T S = −kB T ln ZN S=

(3.40) (3.41)

70

Chapter 3. Thermodynamics and Statistical Physics

Now in Thermodynamics the free energy F is deﬁned by F = U − TS

(3.42)

whence we deduce the relation between the thermodynamical potential F , and the partition function ZN expressing the Statistical Physics properties : F = −kB T ln ZN

(3.43)

Let us write the diﬀerential of the thermodynamical potential F in the case where the inﬁnitesimal work is given by δW = − Xj dxj (see eq. 3.8) : j

dF = dU − T dS − SdT Xj dxj + δQ + µdN − T dS − SdT =−

(3.44)

j

Now δQ = T dS, so that dF = −

Xj dxj − SdT + µdN

(3.45)

j

The interpretation of (3.45) is that, in a speciﬁc inﬁnitesimal reversible process at constant temperature and ﬁxed number of particles, the free energy variation is equal to the work received by the system, whence the name of Helmholtz “free energy,” i.e., available energy, for this thermodynamical potential. For a gas, the inﬁnitesimal work is expressed as a function of pressure, so that dF = −P dΩ − SdT + µdN

(3.46)

The entropy, the pressure, and the chemical potential are obtained as partial derivatives F or of ln ZN : ã ã Å Å ∂(kB T ln ZN ) ∂F = (3.47) S=− ∂T N,Ω ∂T N,Ω ã ã Å Å ∂(kB T ln ZN ) ∂F = (3.48) P =− ∂Ω N,T ∂Ω N,T ã ã Å Å ∂(kB T ln ZN ) ∂F µ= =− (3.49) ∂N Ω,T ∂N Ω,T

3.5.3

System in Contact with Both a Heat Reservoir at Temperature T and a Particle Reservoir

It was shown in 2.5.1 that, during such an exchange of energy and particles, the parameters taking the same value in equilibrium in both the system and

The Thermodynamical Potentials ; the Legendre Transformation

71

the reservoir are kB β = 1/T and kB α = µ/T : two Lagrange parameters thus come into the probabilities, which do not only depend on the energy but also on the number of particles. Thus the probability of occurrence of a macroscopic state of energy Ei and number of particles Ni can be written (see § 2.5.3, eqs. (2.62) and (2.63)) : pi (Ei , Ni ) = so that

pi ln pi =

i

i

=−

1 exp(−βEi + αNi ) ZG

Å ã Ei + αNi − ln ZG pi − kB T

U + αN − ln ZG kB T

(3.50)

(3.51) (3.52)

Consequently, U − αkB N + kB ln ZG T U − T S − αkB T N = −kB T ln ZG S=

(3.53) (3.54)

where N is the average number of particles in the system under study. The thermodynamical potential introduced here is the grand potential A, deﬁned by A = U − T S − µN = F − µN

(3.55)

It is related to the grand canonical partition function through A = −kB T ln ZG

(3.56)

The diﬀerential of A is deduced from its deﬁnition : dA = dF − µdN − N dµ Xj dxj − SdT − N dµ =−

(3.57) (3.58)

j

Like the other thermodynamical potentials, A is extensive. The only extensive variable in A is the one in the inﬁnitesimal work. In the case of a gas, it is the volume so that A = −P Ω

(3.59)

The entropy S and the average number of particles N are obtained by calculating the partial derivatives of A. As a conclusion of this section on the thermodynamical potentials, remember that to a ﬁxed physical situation, corresponding to data on speciﬁc external parameters, is associated a particular thermodynamical potential, which is extremum in equilibrium.

72

3.5.4

Chapter 3. Thermodynamics and Statistical Physics

Transformation of Legendre ; Other Thermodynamical Potentials

In § 3.5.2 and § 3.5.3 we saw two diﬀerent physical situations, associated with two thermodynamical potentials with similar diﬀerentials, except for a change in one variable. The mathematical process that shifts from F (N ) to A(µ) is called “transformation of Legendre.” The general deﬁnition of such a transformation is the following : one goes from a function Φ(x1 , x2 , ...) to a function Γ = Φ(x1 , x2 , ...) − x1 y1 , with y1 =

∂Φ ∂x1

(3.60)

the new function Γ is considered as a function of the new variable y1 . Because of the change of variable and of function, the diﬀerentials of Φ and of Γ are respectively : ∂Φ ∂Φ dx1 + dx2 + . . . ∂x1 ∂x2 ∂Γ ∂Γ dy1 + dx2 + . . . dΓ = ∂y1 dx2

dΦ =

(3.61) (3.62)

Now ∂Γ ∂Γ ∂x1 = · = ∂y1 ∂x1 ∂y1

Å

∂y1 ∂Φ ∂Φ − − x1 ∂x1 ∂x1 ∂x1

ã ·

∂x1 = −x1 ∂y1

(3.63)

Consequently, dΓ = −x1 dy1 +

∂Φ dx2 + . . . ∂x2

(3.64)

The partial derivatives of Γ versus the variables x2 . . . are not modiﬁed with respect to those of Φ for the same variables. In this process one has shifted from the variables (x1 , x2 . . . xn ) to the variables (y1 , x2 . . . xn ). This is indeed the transformation that was performed for example, when shifting from U (S) to F (T ), the other variables Ω, N being conserved, or when shifting from F (N ) to A(µ), the variables Ω and T being unchanged. Other thermodynamical potentials can be deﬁned in a similar way, a speciﬁc potential corresponding to each physical situation. For example, a frequent situation in Chemistry is that of experiments performed at constant temperature and constant pressure (isothermal and isobaric condition) : the exchanges occur between the studied system A and the larger system A which acts as

The Thermodynamical Potentials ; the Legendre Transformation

73

an energy and volume reservoir. We saw that the equilibrium condition, expressing the maximum of entropy of the combined system with respect to exchanges of energy and volume, corresponds to the equality of the temperatures and the pressures in both systems. The corresponding thermodynamical potential is the Gibbs free enthalpy G, in which the number of particles is ﬁxed : G = U − T S + P Ω dG = ΩdP − SdT + µdN

(3.65)

This potential is deﬁned from the free energy F F = U − TS dF = −P dΩ − SdT + µdN

(3.66)

using a transformation of Legendre on the set of variables (volume Ω and pressure P ), that is G = F + P Ω

(3.67)

The chemical potential is interpreted here as the variation of the free enthalpy G when one particle is added to the system, at constant pressure and temperature. The free enthalpy, extensive like all thermodynamical potentials, depends of a single extensive variable, so that G = µN

(3.68)

The chemical potential µ is here the free enthalpy per particle. The free enthalpy is related to the partition function which includes all the possible values of energy and volume of the considered system (see § 2.6) : ZT,P =

e−βEi − γΩi

(3.69)

i

Note 1 : each time a transformation of Legendre is performed, an extensive variable is replaced by an intensive one. Now a thermodynamical potential has to be extensive, so that it should retain at least one extensive variable. Consequently, there is a limit to the number of such tranfomations yielding a new thermodynamical potential. Note 2 : Appendix 4.1 presents a practical application of thermodynamical potentials to the study of chemical reactions.

Summary of Chapter 3 We have expressed the laws of Thermodynamics in the language of Statistical Physics. – Zeroth law : For two systems A and A in thermal equilibrium ã Å 1 ∂S 1 = kB β = = kB β = T ∂E N,Ω T – First law : The diﬀerential of the average energy of a system : dU = d pi Ei = pi dEi + Ei dpi i

i

i

is split into two terms. The inﬁnitesimal work is related to the modiﬁcation of the energy states of the system due to the change of external parameters (volume, etc.) δW = pi dEi i

The inﬁnitesimal heat exchanged δQ is associated with variations of probabilities of occurrence of the microstates Ei dpi δQ = i

– Second law : We have identiﬁed the thermodynamical entropy with the statistical entropy as introduced in the ﬁrst chapters of this book and have shown that 1 kB β = . T 75

76

Summary of Chapter 3

In Statistical Physics, an irreversible process corresponds to an increase of the number of accessible microstates for the considered system, and consequently to an increase of disorder and of entropy. – Third law : The entropy per particle S/N vanishes when the temperature tends to the absolute zero. – The adapted thermodynamical potential depends on the considered physical situation, it is extremum in equilibrium. It is related to the microscopic parameters of Statistical Physics through the partition function ; its partial derivatives give access to the measurable physical parameters. Using a transformation of Legendre, one gets a new thermodynamical potential, with one new variable (see the table next page).

Parameters

Thermodynamical

dS =

P dU µ dΩ + − dN T T T 1 ∂S ... = T ∂U Ω,N

S = kB ln W

Thermodynamical Potential

microstates

S=−

∂F ∂T

e−βEn

Ω,N

...

dF = −P dΩ − SdT + µdN

F = −kB T ln ZN F = U − TS

ZN (β) =

W microscopic conﬁgurations at E

Statistical Description

1 kB T

exchange of energy with a heat reservoir at ﬁxed N

β=

isolated system E to δE

Physical Conditions

Canonical

Lagrange Parameter(s)

Microcanonical

Ensemble

eαN ZN (β)

N = −

Å

∂A ∂µ

Ω,T

ã

(average value)

dA = −P dΩ − SdT − N dµ

A = −kB T ln ZG A = U − T S − µN

N

1 µ , α = βµ = kB T kB T

ZG (α, β) =

β=

exchange of energy with a heat reservoir and of particles with a particle reservoir

Grand Canonical

Summary of Chapter 3 77

Chapter 4

The Ideal Gas 4.1

Introduction

After having stated the general methods in Statistical Physics and bridged its concepts with your previous background in Thermodynamics, we are now going to apply our know-how to the molecules of a gas, with no mutual interaction. The properties of such a gas, called “ideal gas” (“perfect gas” in French) are discussed in Thermodynamics courses. Here the contribution of Statistical Physics will be evidenced : starting from a microscopic description in Classical Mechanics, it allows to explain the observed macroscopic properties. A few notions of kinetic theory of gases will ﬁrst be introduced in § 4.2 and § 4.3 we will discuss to what extent classical Statistical Physics, as used up to this point in this book, can apply to free particles. This chapter is “doubly classical” since, like in the three preceding chapters, Classical Statistics is used, but now for particles with an evolution described by Classical Mechanics : both the states themselves and their occurrence are classical. The canonical ensemble is chosen in § 4.4 to solve the statistical properties of the ideal gas and physical consequences are deduced ; in the Appendices, all that we learnt in this chapter and the preceding ones is applied in a practical problem : the interpretation of chemical reactions in gas phase is done in Thermodynamics (Appendix 4.1) and in Statistical Physics (Appendix 4.2).

79

80

4.2

Chapter 4. The Ideal Gas

Kinetic Approach

This section, which has no relation with Quantum Mechanics or Statistical Physics in equilibrium, will allow us to obtain insight into the simplest physical parameters describing the motion of molecules in a gas, and to introduce some orders of magnitude. If necessary, we will refer to the Thermodynamics courses on the ideal gas.

4.2.1

Scattering Cross Section, Mean Free Path

Consider a can ﬁlled with a gas, in the standard conditions of temperature and pressure (T = 273 K, P = 1 atmosphere). Each molecule has a radius r, deﬁned by the average extension of the valence orbitals, i.e., a fraction of a nanometer. If in its motion a molecule comes too close to another one, there will be a collision. One deﬁnes the scattering cross section which is the circle of radius r around a molecule, inside which no other molecule can penetrate.

l=

ντ

r

d Fig. 4.1: Deﬁnitions of the scattering cross section and of the mean free path. The mean distance between two collisions l, or mean free path, is deduced from the mean volume per particle Ω/N and the scattering cross section by l(πr2 ) =

Ω = d3 N

(4.1)

where d is the average distance between particles (Fig. 4.1). From l, one deduces the mean time τ between two collisions, or collision time, by writing l = ντ where ν is the mean velocity at temperature T , of the order of the justiﬁcation in § 4.4.2).

(4.2) kB T /m (see

Kinetic Approach

81

For nitrogen molecules with r = 0.42 nm, at 273 K and under a pressure of one atmosphere one ﬁnds l ≈ 70 nm, ν ≈ 300 m.sec−1 , τ ≈ 2 × 10−10 sec. The length l is to be compared with the average distance between particles d = 3.3 nm. These values show that in the gas collisions are very frequent at the time scale of a measure, which ensures that equilibrium is rapidly reached after a perturbation of the gas ; at the molecular scale distances covered between collisions are very large (of the order of a hundred times the size of a molecule), thus the motion of each molecule mostly occurs as if it was alone, that is, free.

4.2.2

Kinetic Calculation of the Pressure

Let us now recall the calculation, from kinetic arguments, of the pressure of a gas, i.e., of the force exerted on a wall of unit surface by the molecules. The aim is to evaluate the total momentum transferred to this wall by the molecules impinging on it during the time interval ∆t : indeed in this chapter Classical Mechanics does apply, so that the total force on the wall is given by the Fundamental Equation of Dynamics : Ftotal =

∆ pi i

(4.3)

∆t

Here the momentum brought by particle i is ∆ pi . The resulting force is normal since the wall is at rest, it is related to the pressure P and to the considered wall surface S through |Ftotal | = P S

(4.4)

One will now evaluate the total momentum transferred to the surface, which is the diﬀerence between the total momentum of the particles which reach the wall and that of the particles which leave it (Fig. 4.2). pi f (r, p i , t) The incident particles, with pz > 0, bring a total momentum i

to the wall, chosen as xOy plane. The probability density in equilibrium f (r, pi , t) is assumed to be stationary, homogeneous, and isotropic in space (hypotheses stated by James Clerk Maxwell in 1859). It thus reduces to an even function f ( p) versus the three components of the momentum and satisﬁes to (4.5) f ( p)d3 p = 1 Consider the particles, of momentum equal to p to d3 p, incident on the surface S during the time interval ∆t. They are inside an oblique cylinder along the

82

Chapter 4. The Ideal Gas

pz ∆t m S

p

O

z

p

Fig. 4.2 : Collisions of molecules on a wall. Spz pz ∆t and volume ∆t (the direction z is m m chosen along the normal to the surface). For a particles density ρ = N/Ω, the number of incident particles is

direction of p, of basis S, height

Spz ∆t ρf ( p)d3 p m

(4.6)

Each of them transfers its momentum p to the wall. The total momentum brought by the incident particles is thus equal to S ∆ pi = ∆tρ pz pf ( p)d3 p (4.7) m p >0 z inc This sum is restricted to the particles traveling toward the surface, of momentum pz > 0. A similar calculation leads to the total momentum of the particles which leave the surface, or in other words “were reﬂected,” by the wall. The volume (−pz ) is S ∆t, the total momentum taken from the wall is equal to m S ∆ pi = ∆tρ (−pz ) pf ( p)d3 p (4.8) m pz 0. Indeed one will always perform an integration by parts of a seconddegree term in energy in the numerator to express the corresponding factor of the denominator partition function, which will provide the factor kB T /2. But this property is not valid for a polynomial of degree diﬀerent than two, for example for an anharmonic term in x3i . This result constitutes the “theorem of energy equipartition”, established by James C. Maxwell as early as 1860 : For a system in canonical equilibrium at temperature T , the average value of the energy per particle and per quadratic degree of 1 freedom is equal to kB T . 2 We have already used this property several times, in particular in § 4.3 to ﬁnd the root-mean-square momentum of the ideal gas at temperature T . It is also useful to determine the thermal ﬂuctuations of an oscillation, providing 1 1 a potential energy in kx2 for a spring, in Cϑ2 for a torsion pendulum, 2 2 and more generally to calculate the average value at temperature T of a hamiltonian of the form h= ai x2i + bj p2j (4.33) i

j

where ai is independent of xi , bj does not depend of pj , and all the coeﬃcients ai and bj are positive (which allows to use this property for an energy development around a stable equilibrium state). A case of application of the energy equipartition theorem is the speciﬁc heat at constant volume associated to U . We recall its expression ã Å ∂U Cv = (4.34) ∂T Ω,N

Classical Statistics Treatment of the Ideal Gas

89

In the case of a monoatomic ideal gas, Cv only contains the contribution of 3 the three translation degrees of freedom and is thus equal to N kB . 2

4.4.3

Free Energy ; Physical Parameters (P, S, µ)

We saw in chapter 3 that the canonical partition function Zc is related to the free energy through F = U − T S = −kB T ln Zc

(4.35)

In the case of an ideal gas, the diﬀerential of F is dF = −P dΩ − SdT + µdN

(4.36)

The volume Ω and the number of particles N are extensive variables, the temperature T is intensive. The chemical potential, already introduced in § 3.5.3, is here the variation of free energy of the system at ﬁxed temperature and volume when one particle is added. Substituting the expression of Zc obtained in § 4.4.1, one obtains : Å ã Ω F = −kB T N ln 3 (2πmkB T )3/2 + ln CN h ã Å Ω 1 3/2 F = −N kB T ln ln C (2πmk T ) + ln N + B N N h3 N

(4.37) (4.38)

For the potential F to be extensive, the quantity between brackets must be Ω should appear in the ﬁrst term and intensive. This dictates that the ratio N that the remaining terms should be constant : ln N +

1 ln CN = constant N

(4.39)

i.e., ln CN = −N ln N + constant × N

(4.40)

From now on, we will take CN =

1 , i.e. ln CN = −N ln N + N + O(ln N ) N!

(4.41)

so that the constant in Eq. (4.39) is equal to unity. This choice will be justiﬁed in § 4.4.4 by a physical argument, the Gibbs paradox, and will be demonstrated in § 6.7. Then ï Å Å ã ò ã Ω eΩ 3/2 (2πmk T ) T ln +1 = −N k (4.42) F = −N kB T ln B B N h3 N λ3th

90

Chapter 4. The Ideal Gas

One deduces the pressure, the entropy and the chemical potential of the assembly of N particles : Å ã ∂F N kB T (4.43) = P =− ∂Ω N,T Ω ã ã Å Å 5 ∂F Ω + N kB S=− = N kB ln (4.44) ∂T N,Ω N λ3th 2 ã ã Å Å Ω ∂F (4.45) = −kB T ln µ= ∂N T,Ω N λ3th One identiﬁes the ﬁrst of these relations as the ideal gas law, that we now just re-demonstrated using Statistical Physics. The expression obtained for S(Ω, N, T ) is extensive but does not satisfy the Nernst principle, since it does not vanish when the temperature tends to zero : the entropy would become very negative if the de Broglie thermal wavelength was becoming too large with respect to the mean distance between particles. This conﬁrms what we already expressed in § 4.3, that is, the ideal gas model is no longer valid at very low temperature. In the validity limit of the ideal gas model, λth d is equivalent to Ω/N λ3th very large, thus to µ very negative (µ → −∞). Remember that the measure of the classical N -particle phase space must be N 1 d3ri d3 pi taken equal to to ensure the extensivity of the thermodynaN ! i=1 h3 mical potentials F and S. This will be justiﬁed in chapter 6 : the factor CN expresses the common classical limit of both Quantum Statistics.

4.4.4

Gibbs Paradox

At the end of the 19th century, John Willard Gibbs used the following physical argument to deduce the requirement of the constant CN = N1 ! in the expression of the inﬁnitesimal volume of the phase space : consider two containers of volumes Ω1 and Ω2 , containing under the same density and at the same temperature T , respectively, N1 and N2 molecules of gas (Fig. 4.3). One then opens the partition which separates the two containers, the gases mix and the ﬁnal total entropy is compared to the initial total one. If the gases in the two containers were of diﬀerent chemical natures, the disorder and thus the entropy increase. On the other hand, if both gases are of the same nature and the molecules indistinguishable, the entropy is unchanged in the process. The detailed calculation of the diﬀerent terms in the total entropy, before

Conclusion

91

N1 et N2 of the same nature Ω 1 + Ω2 N1 + N2 T Ω1 N1 S1 T

Ω2 N2 S2 T

N1 N2 = Ω1 Ω2

S = S1 + S2 N1 et N2 of diﬀerent natures Ω 1 + Ω2 N1 , N2 T S > S1 + S2

Fig. 4.3: Mixture of two gases at the same temperature and with the same density. and after opening the partition, is simple and uses the partial derivative with respect of temperature of expression (4.38) of the free energy, including the constant CN : if one assumes that CN = 1, there will be a variation of the total entropy even for a mixture of molecules of the same nature. When one introduces the factor CN = 1/N !, the result corresponds to the physical prediction. The choice of this factor is not surprising once one considers that molecules of the same chemical nature are indistinguishable : for a system of N indistinguishable particles, there are N ! possible permutations of these particles among their one-particle states, providing the same macroscopic state. In order to account only once for this state in the enumeration that will lead to the physical predictions, among others of the entropy, it is natural to choose the value CN = 1/N !.

4.5

Conclusion

Using Statistical Physics, we have been able to ﬁnd again the well-known results of Thermodynamics on the ideal gas and to justify them. It seems interesting, now that we come to the end of the study of Classical Statistics, to present the respective contributions of Thermodynamics (Appendix 4.1) and of Statistical Physics (Appendix 4.2) on a practical example. This will be done on the analysis of the chemical reactions. In the Statistical Physics part, one will show in details the various microscopic terms of the hamiltonian of a gas of diatomic molecules ; this allows one to

92

Chapter 4. The Ideal Gas

understand the origin of a law very commonly used in Chemistry, that is, the law of mass action. Enjoy your reading ! Reading these two Appendixes is not a prerequisite to understanding the next chapters.

Summary of Chapter 4 First we have presented some orders of magnitude, relative to nitrogen molecules in standard conditions of temperature and pressure. They justify that the hypotheses of the ideal gas, of molecules with no mutual interaction, is reasonable : indeed the distances traveled by a molecule between collisions are very large with respect to a molecule radius and the characteristic time to reach equilibrium is very short as compared to a typical time of measurement. Then the kinetic calculation of a gas pressure was recalled. The domain of validity of the ideal gas model, in the framework of both Classical Mechanics and Classical Statistics, is the range of high temperatures and low densities : the criterion to be fulﬁlled is Å ã1/3 Å ã1/3 Ω h Ω

d= 2πmkB T , or λth = √ h N N 2πmkB T where λth is the thermal de Broglie wavelength and d the mean distance between particles. The statistical treatment of the ideal gas is done in this chapter in the canonical ensemble, the N -particle canonical partition function is given by ïÅ ã òN Ω 3/2 Zc = CN (2πmk T ) B h3 To ensure that the free energy and the entropy are extensive and to solve the Gibbs paradox, one must take CN = 1/N !. The average value at temperature T of the energy associated to a quadratic 1 degree of freedom is equal to kB T (theorem of the equipartition of energy). 2 This applies in particular to the kinetic energy terms for each coordinate. The chemical potential of the ideal gas is very negative ; its entropy does not satisfy to the third law of Thermodynamics (but the ideal gas model is no longer valid in the limit of very low temperatures). 93

Appendix 4.1 Thermodynamics and Chemical Reactions We will consider here macroscopic systems composed of several types of molecules or chemical species, possibly in several phases : one can study a chemical reaction between gases, or between gases and solids, etc. The ﬁrst question that will be raised is “What is the spontaneous evolution of such systems ?” ; then the conditions of chemical equilibrium will be expressed. These types of analyses, only based on Thermodynamics, were already introduced in your previous studies. Now Appendix 4.2 will evidence the contribution of Statistical Physics. 1. Spontaneous Evolution In the case of a thermally isolated system, the second law of Thermodynamics indicates that a spontaneous evolution will take place in a direction corresponding to an increase in entropy S, S being maximum in equilibrium.

Q A A T0

Atot Fig. 4.4: The system A is in thermal contact with the heat reservoir A , with which it exchanges the heat Q. The combined system is isolated. 95

96

Appendix 4.1

Chemical reactions are mostly done at constant temperature : one then has to consider the case of a system A in thermal contact with a heat reservoir A at temperature T0 (Fig. 4.4). The combined system Atot , made of A and A is an isolated system, the entropy Stot of which increases in a spontaneous process : ∆Stot ≥ 0

(4.46)

The entropy variation of the combined system ∆Stot = ∆S + ∆S

(4.47)

can be expressed as a function of parameters of the system A alone. Indeed, if the amount of heat Q is given to A by the heat reservoir, the reservoir absorbs −Q and its entropy variation is equal to ∆S = −Q/T0 : the transfer of Q does not produce a variation of T0 and one can imagine a reversible sequence of isothermal transformations at T0 producing this heat exchange. Besides, the internal energy variation in A, i.e., ∆U , is, from the ﬁrst law of Thermodynamics, ∆U = W + Q

(4.48)

where W is the work operated in the considered transformation. Consequently, ∆Stot = ∆S −

(∆U − W ) −∆F0 + W = T0 T0

(4.49)

where, by deﬁnition, F0 = U − T0 S, which is identical to the Helmholtz free energy F = U − T S of system A if the temperature of A is that of the heat reservoir (in an oﬀ-equilibrium situation T diﬀers from T0 and F is not identical to F ). From the condition of spontaneous evolution one deduces that −∆F0 ≥ (−W )

(4.50)

which expresses that the maximum work that may be delivered by A in a reversible process is (−∆F0 ), whence the name of “free energy.” If the volume of system A is maintained constant and there is no other type of work, W = 0, so that the condition for spontaneous evolution is then given by ∆F0 ≤ 0

(4.51)

Appendix 4.1

97

This condition for a system in contact with a heat reservoir replaces the condition ∆Stot ≥ 0 for an isolated system. A more realistic situation is the one in which system A is kept at constant temperature and constant pressure. This is the case in laboratory experiments when a chemical reaction is performed at ﬁxed temperature T0 and under the atmospheric pressure P0 : system A is exchanging heat and work with a very large reservoir A , which can thus give energy to A without any modiﬁcation of T0 and also exchange a volume ∆Ω with A without any change of its pressure P0 (Fig. 4.5). A T0 P0

A Q T ∆Ω P

Atot Fig. 4.5: System A is in contact with the larger system A , with which it is exchanging the amount of heat Q and the volume ∆Ω. Similarly, one considers that the combined system made of A and A is isolated, so that in a spontaneous process ∆Stot = ∆S + ∆S ≥ 0

(4.52)

If A absorbs Q from A , ∆S = −Q/T0 , where T0 is again the temperature of A . The variation of internal energy of A includes the received heat Q, the mechanical work −P0 ∆Ω received by A from A , and also W ∗ , work of any other origin (electrical for example) that A may have received : ∆U = Q − P0 ∆Ω + W ∗

(4.53)

Consequently, ∆Stot = ∆S − i.e.,

∆Stot =

Q T0 ∆S − (∆U + P0 ∆Ω − W ∗ ) = T0 T0

−∆G0 + W ∗ T0

(4.54) (4.55)

98

Appendix 4.1

Here we deﬁned G0 = U + P0 Ω − T0 S, which identiﬁes to the Gibbs free enthalpy of system A, that is, G = U + P Ω − T S, when P = P0 and T = T0 . A spontaneous evolution will be expressed through the condition −∆G0 + W ∗ ≥ 0

(4.56)

When all the external parameters of system A, except its volume, are kept ﬁxed, W ∗ is null and the condition of spontaneous evolution is then ∆G0 ≤ 0

(4.57)

Consequently, when a system is in contact with a reservoir so that it remains at ﬁxed temperature and pressure and can exchange work only with this reservoir, the stable equilibrium is characterized by the condition of minimum of G0 . 2. Chemical Equilibrium, Law of Mass Action : Thermodynamical Approach Consider a homogeneous system (for example in gaseous phase) with several types of molecules : here we will study the dissociation reaction of iodine vapor at temperatures in the 1000 K range : I2 2I In this reaction the number of iodine atoms is conserved between the species corresponding to the two members of the chemical equation, which can also be expressed in the form 2I − I2 = 0 the species disappearing in the reaction carrying a negative sign, the one appearing a positive one. In the same way, the most general chemical equilibrium will be written n

bi Ni = 0

(4.58)

i=1

The bi are the integral coeﬃcients of the balanced reaction for the Bi molecules. The number Ni of molecules of Bi type in the system is changed if the equilibrium is displaced but the total number of atoms of each species in the system is conserved. The variations of the numbers Ni are thus proportional to the coeﬃcients bi of the chemical equilibrium dNi = bi dλ

for any i

where dλ is a proportionality constant. dNi is positive for the molecules appearing, negative for those vanishing in the reaction.

Appendix 4.1

99

Consider a chemical equilibrium at ﬁxed temperature T and pressure P , thus characterized by dG(T, P ) = 0. In such conditions the equilibrium is expressed by : dG = At ﬁxed T and P, dG =

n i=1 n

dGi =

n

(−Si dT + Ωi dP + µi dNi )

i=1

(4.59)

µi dNi = 0

i=1

Å

ã ∂G Here µi = is the chemical potential of species Bi . Replacing dNi ∂Ni T,P by its expression versus λ, one obtains the relation characterizing the equilibrium : n

bi µi = 0

(4.60)

i=1

The chemical potentials depend on the experimental conditions, in particular on the concentrations of the diﬀerent species. Assume that these reagents are ideal gases and look for the dependence of µi on the partial pressure Pi of one of these gases. Since Gi is extensive and depends on a single extensive parameter Ni , the number of molecules of the species i, then one must have Gi = µi Ni

(4.61)

The dependence in temperature and pressure of µi is thus that of Gi . For an ideal gas, at ﬁxed temperature, Gi = Ui + Pi Ω − T Si

(4.62)

introducing the partial pressure Pi of species i, such that Pi Ω = Ni kB T . Consequently, dGi = −Si dT + ΩdPi + µi dNi Å

∂Gi ∂Pi

ã

Å = Ni

T,Ni

∂µi ∂Pi

ã =Ω= T,Ni

Ni kB T Pi

(4.63)

(4.64)

i.e., by integrating (4.64), µi = µ0i (T ) + kB T ln

Pi P0

(4.65)

100

Appendix 4.1

where µ0i (T ) is the value of the chemical potential of species i at temperature T and under the standard pressure P0 . The general condition of chemical equilibrium is thus given by Å ãbi n n Pi bi µ0i + kB T ln =0 P0 i=1 i=1

(4.66)

One writes ∆G0m = bi µ0i N i=1 n

(4.67)

∆G0m , the molar free enthalpy at the temperature T of the reaction, is the diﬀerence between the molar free enthalpy of the synthesized products and that of the used reagents, N is the Avogadro number. One also deﬁnes Kp =

ã n Å Pi bi i=1

P0

(4.68)

For example, in the dissociation equilibrium of iodine molecules, Kp =

(PI /P0 )2 (PI2 )/P0

(4.69)

One recognizes in Kp the constant of mass action. The above equation gives 0

Kp = e−∆Gm /RT

(4.70)

where R = N kB is the ideal gases constant, R = 8.31 J.K−1 . From the tables of free enthalpy values deduced from experiments, one can deduce the value of Kp . From the experimental study of a chemical equilibrium at temperature T , one can also deduce the value of Kp , and thus of ∆G0m , at temperature T : if the temperature of the experiment is modiﬁed from T to T , the partial pressures are changed, Kp (T ) becomes Kp (T ), such that ln Kp (T ) = −∆G0m (T )/RT

(4.71)

By deﬁnition, at a given T , ∆G0m = ∆H 0 − T ∆S 0 , where ∆H 0 is the molar enthalpy of the reaction and ∆S 0 its entropy variation. If T and T are close enough, one makes the approximation that neither ∆H 0 nor ∆S 0 noticeably vary between these temperatures. Then Å ã 1 1 (4.72) − ln Kp (T ) − ln Kp (T ) = −∆H 0 RT RT

Appendix 4.1

101

i.e., also ∆H 0 d ln Kp (T ) = dT RT 2

(4.73)

From the enthalpy ∆H 0 , one can predict the direction of displacement of the equilibrium when the temperature increases : if ∆H 0 > 0, that is, the enthalpy increases in the reaction and the reaction is endothermal, the equilibrium is thus displaced toward the right side when T increases. This is in agreement with the principle of Le Châtelier : If a system is in stable equilibrium, any spontaneous change of its parameters must lead to processes which tend to restore the equilibrium. As an application, the experimental data below on the dissociation of iodine allow to deduce the enthalpy of this reaction in the vicinity of 1000 K. For a ﬁxed total volume of 750.0 cm3 , in which a total number n of molecules I2 was introduced at temperature T , the equilibrium pressure is P : T (K) P (10−2 atm) n(10−4 moles)

973 6.24 5.41

1 073 7.50 5.38

1 173 9.18 5.33

As an exercise, one can calculate Kp at each temperature and verify that ∆H 0 = 157 kJ.mole−1 . This reaction is endothermal. The example presented here suggests the rich domain, but also the limitations of the applications of Thermodynamics in Chemistry : Thermodynamics does not reduce to a game between partial derivatives, as some students may believe ! It allows one to predict the evolution of new reactions, from tables deduced from measures. However, only the contribution of Statistical Physics allows us to predict these evolutions from the sole spectroscopic data and microscopic models. It is this latter approach that will be followed now. Thanks to the understanding acquired in this ﬁrst part of the Statistical Physics course, we will be able to deduce the constant of mass action from the quantum properties of the chemical species of the reaction.

Appendix 4.2 Statistical Physics and Chemical Reactions The same example of dissociation of the iodine molecule at 1000 K, in gas phase, will now be treated by applying the general methods of Statistical Physics. The calculations will be completely developed, which may be regarded as tedious but yet they are instructive ! 1. Energy States : The Quantum Mechanics Problem Iodine I is a monoatomic gas, its only degrees of freedom are those of translation. The electronic structure of its fundamental state is 2 P3/2 (notation 2S+1 LJ : orbital state = 1 whence the notation P ; spin 1/2, i.e., 2S + 1 = 2 ; total kinetic moment equal to J = 3/2, therefore the degeneracy 4). Iodine has a nuclear moment J=5/2. The hamiltonian of the iodine atom in the gas phase is therefore given by HI = HI transl + HI elec + HI nucl

(4.74)

The molecule I2 has six degrees of freedom, which can be decomposed into three degrees of translation of its center of mass, two of rotation and one of vibration : HI2 = HI2 transl + Hrot + Hvibr + HI2 elec + HI2 nucl

(4.75)

Since this molecule is symmetrical, one should account for the fact that the two iodine atoms are indistinguishable inside the molecule and thus the physics of the system is invariant in a rotation of 180 degrees around an axis perpendicular to the bond in its middle. The fundamental electronic state of molecular iodine is not degenerate. It lies lower than the fundamental electronic state of atomic iodine. The dissociation energy, necessary at zero temperature to transform a molecule into two iodine 103

104

Appendix 4.2

atoms, is equal to Ed = 1.542 eV, which corresponds to 17,900 K : one should bring Ed /2 per atom to move from the state I2 to the state I. The nuclear spin of iodine has the value J = 5/2, both under the atomic and molecular forms ; one assumes that no magnetic ﬁeld is applied. The splittings between energy levels of I or I2 are known from spectroscopic measures : Raman or infrared techniques give access to the distances between rotation or vibration levels ; the distance between electronic levels are generally in the visible (from 1.5 eV to 3 eV) or the ultraviolet range. One now expresses the eigenvalues of the diﬀerent terms of HI and HI2 . a. Translation States One has to ﬁnd the hamiltonian eigenstates : in the case of the iodine atom of mass mI , Htr =

pˆ2 + V (r) 2mI

(4.76)

where V (r) is the conﬁnement potential limiting the particle presence to a rectangular box of volume Ω = Lx Ly Lz . √ The eigenstates are plane waves ψ = (exp ik · r)/ Ω ; the k values are quantized and for the Born-Von Kármán periodic limit conditions [ψ(x+Lx ) = ψ(x)] (see § 6.4.2) the k components satisfy : kx Lx = nx 2π ; ky Ly = ny 2π ; kz Lz = nz 2π with nx , ny , nz positive, negative or null integers. The corresponding energies are : 2 2 k 2mI ñÅ ã2 Å ã2 Å ã2 ô nx h2 ny nz = + + 2mI Lx Ly Lz

ε I tr =

(4.77)

The translation energies εI2 tr of the molecule I2 are obtained by replacing mI by mI2 in the above expression of εItransl . All these states are extremely close : if Ω is of the order of 1 cm3 and the dimensions of the order of 1 cm, the energy of the state nx = 1, ny = nz = 0 is given by, for I (mI = 127 g) Å ã2 a0 1 mH × (0.53 × 10−8 )2 = 3 × 10−18 eV ! · = 13.6 × 13.6 eV × mI Lx 127

Appendix 4.2

105

The distances between energy levels are in the same range, thus very small with respect to kB T , for T = 1000K. This discussion implies that the ideal gas model is valid at room temperature for I and I2 (still more valid than for N2 , since these molecules are heavier). The partition functions for their translation states will take the form (4.26)(4.27).

b. Rotation States The iodine molecule is linear (Fig. 4.6), its inertia momentum around an Å ã2 Re axis normal to the bond is I = 2m = µRe2 with µ = mI /2 and Re 2 the equilibrium distance between the two nuclei. Since mI = (127/N )g this momentum is particularly large at the atomic scale. The rotation hamiltonian is 1 ˆ2 J 2I and shows that the problem is isotropic in the center of mass frame. Hrot =

(4.78)

Re

mI

mI

Fig. 4.6 : Structure of the I2 molecule. Its eigenvalues are : εrot (J) =

2 J(J + 1) = hcBJ(J + 1), 2I

where

B=

4πcI

is given in cm−1 (4.79)

(one recalls the deﬁnition hν = hc/λ = hcσ ; the unit cm−1 for σ is particularly adapted to infrared spectroscopy). The selection rules ∆J = ±1 for allowed electric dipolar transitions between rotation levels correspond to εrot (J + 1) − εrot (J) = hc · 2B(J + 1). For iodine B = 0.0373 cm−1 , 2hcB/kB = 0.107 K, to be compared to HC where B = 10.6 cm−1 , 2hcB/kB = 30.5 K. The rotation levels are thus still very close at the thermal energy scale.

106

Appendix 4.2

c. Vibration States In a ﬁrst approximation the vibration hamiltonian of the molecule I2 can be described as a harmonic oscillator hamiltonian, with a parabolic potential well centered at the equilibrium distance Re between the two nuclei, that is, V (R) = 12 K(R−Re )2 (Fig. 4.7). (In this ﬁgure a realistic potential is sketched in full line ; the dotted line represent its approximated harmonic potential.) ε

3 2 1 n=0 Re

R

Fig. 4.7 : Potential well associated to the vibration of the molecule I2 .

Hvibr =

1 pˆ2R + K(R − Re )2 2µ 2

(4.80)

∂ with pˆR = −i , µ being the reduced mass of the system. In this approxi∂R mation, the vibration energies are given by ã Å K 1 εvibr (n) = n + ω with ω = (4.81) 2 µ The transitions ∆ε = ω = hcσv are in the infrared domain. For I2 σv = 214.36 cm−1 , which corresponds to λ = 46.7µm or hcσv /kB = 308 K (for F2 the corresponding temperature is 1280 K).

d. Electronic States The splitting between fundamental and excited states, for both I and I2 , is of the order of a few eV, corresponding to more than 10,000 K. The origin of the electronic energies (Fig. 4.8) is taken at the I2 fundamental state, so that the energy of the I fundamental state is EI , with 2EI = Ed = 1.542 eV, where Ed is the dissociation energy.

Appendix 4.2

107

εe2 εe1 εe1 g=4

εe0 EI

g=1 I2

εe0

I

Fig. 4.8 : Electronic states of I2 and I. e. Iodine Nuclear Spin It is J = 5/2 on both species I and I2 . 2. Description of the Statistical State of the System : Calculation of the Partition Functions It was stated in Ch.2 that for a macroscopic system the predictions for the experimental measurements of physical parameters were the same whatever the chosen statistical ensemble. Here we will work in the canonical ensemble, where the calculations are simpler. At temperature T , the number of iodine molecules is NI2 , the number of atoms produced by the dissociation of the molecules NI . The partition function for a single iodine molecule is written zI2 , that of an isolated atom zI . Because of the particles indistinguishability, the total canonical partition function Zc is given by : Zc =

1 1 (zI )NI2 (zI )NI NI2 ! 2 NI !

(4.82)

It is related to the free energy F of the system F = −kB T ln Zc

(4.83)

the partial derivatives of which, with respect to the numbers of particles, are the chemical potentials : ã Å Å ã ∂F zI2 zI µI = −kB T ln (4.84) = −kB T ln µI2 = ∂NI2 NI2 NI Since the total energy of an iodine atom or molecule is a sum of contributions, the partition function of a corresponding atom or a molecule is a product of

108

Appendix 4.2

factors : thus for an I2 molecule εI2 = εI2 tr + εrot (J) + εvibr (n) + εel + εnucl

(4.85)

zI2 = zI2 tr · zrot · zvibr · zel · znucl

(4.86)

with :

2 1 3 r d3 p e−βp /2mI2 d 3 h = gJ e−βεJ = (2J + 1)e−βhcBJ(J+1)

zI2 tr = zrot

J

zvibr =

(4.88)

J

gv e−βεv =

v

zel =

(4.87)

1

e−βω(n+ 2 )

(4.89)

n

ge e−βεe

(4.90)

e

znucl = (2J + 1)2 for I2 ,

znucl = 2J + 1 for I

(4.91)

The factors gr , gv , ge express the degeneracies of the rotation, vibration and electronic energy levels. The value of the nuclear spin is J , there are two nuclei in I2 . One will ﬁrst evaluate the various partition functions, noting that a large value of the partition function is related to a large number of accessible states. The examined chemical potentials will be sums of terms, associated with the zI zI diﬀerent z factors. Note that one has to calculate 2 (or ) (see (4.84)), NI2 NI thus to divide a single factor of zI2 by NI2 : we will associate the factor NI2 to the sole translation factor, then it no longer comes into the partition functions for the other degrees of freedom. (The indiscernability factor is thus included in the translation degree of freedom.) Then : ln

ztr I2 zI2 = ln + ln zrot + ln zvibr + ln zel + ln znucl NI2 NI2

(4.92)

a. Translation The calculation of ztr I was already done for the ideal gas : Å

ztr I

2πmI kB T =Ω· h2

ã3/2 =

Ω NI kB T 1 = · (λth I )3 Pi (λth I )3

(4.93)

Pi is the partial pressure, such that Pi Ω = NI kB T and λth I is the thermal de Broglie wavelength at T for I. Since we are in the classical framework, ztr I

Appendix 4.2

109

must be very large. Numerically λth , expressed in picometers (10−12 m), is given by 1749 · 10−12 m λth = (TK × mg·mole−1 )1/2 At 1000 K, for the iodine atom mI = 127g,

λth I = 4, 91 10−12 m.

In the same condition, for the I2 molecule,

λth I2 = 3.47 10−12 m.

For a partial pressure equal to the atmospheric pressure and a temperature T = 1000 K, one obtains zI tr = 1.15 × 109 NI zI2 tr = 3.26 × 109 NI2 These numbers are huge !

b. Rotation The factor 2J + 1 is the degeneracy of the energy level hcBJ(J + 1). Besides, the linear molecule I2 is symmetrical, so that after a 180 degree rotation one gets a new state, indistinguishable from the initial one (this would not be the case for HC for example). One thus has to divide the above expression (4.88) by two. This is equivalent to saying that the wave function symmetry allows to keep only half of the quantum states. Since the energy distance between rotation levels is small with respect to kB T , one can replace the discrete sum by an integral 1 1 ∞ (2J + 1)e−βhcBJ(J+1) (2J + 1)e−βhcBJ(J+1) dJ 2 2 0 J (4.94) 1 1 kB T = 2 βhcB 2hcB This approximation implies that zrot is a large number. Numerically, the contribution of zrot , equal to 1/2βhcB, is the ratio 1000 K/0.107 K=9346.

c. Vibration zvibr = e−βω/2

∞ n=0

e−βωn =

e−βω/2 1 − e−βω

(4.95)

At 1000 K, for I2 , the term βω is 0.308 : the quantum regime is valid and this expression cannot be approximated. It has the value 3.23.

110

Appendix 4.2

d. Electronic Term

ze =

ge e−βεe

(4.96)

e

Owing to the large distances between electronic levels, one will only consider the fundamental level, four times degenerated for I, nondegenerated for I2 . At 1000 K, the term zeI is equal to 4 e−(Ed /2kB T ) = 4e−8.94 = 5.25 × 10−4 , whereas for the same choice of energy origin zeI2 = 1.

3. Prediction of the Action of Mass Constant Let us return to the deﬁnition relations (4.68) to (4.70) of Appendix 4.1, Kp = exp(−∆G0m /RT ) = exp(−

n

bi µ0i /kB T )

(4.97)

i=1

where the µ0i are the chemical potentials under the atmospheric pressure. For our example : Kp = exp[−(2µ0I − µ0I2 )/kB T ]

(4.98)

The chemical potentials are expressed versus the partition functions calculated above : Å 0 Å 0 ã ã 1 z I tr z I2 tr 0 0 (2µI − µI2 ) = −2[ln + ln zeI ] + ln + ln zrot + ln zvib kB T NI NI2 (4.99) Here the translation partition functions are calculated under the atmospheric pressure (standard pressure) ; the terms associated to the nuclear spins 2 ln(2J + 1) exactly compensate and ln ze I2 = 0. From the data of the previous section one obtains : z 0 I tr = 20.87 , NI z 0 I2 tr ln = 21.91 , NI2 ln

ln zeI = −7.56 ln zrot = 9.14 , ln zvibr = 1.17

Kp = e−5.6 3.7 × 10−3

Appendix 4.2

111

One can also predict, from the calculation of the temperature variation of Kp , a value for the enthalpy of the iodine dissociation reaction : ∆H d ln Kp = dT RT 2

(4.100)

i.e., d dT

Ç

−2µ0I + µ0I2 kB T

å =

∆H RT 2

(4.101)

∆H is thus deduced from the logarithmic derivatives versus T , at constant P , of the diﬀerent partition functions : the translation partition functions vary ã Å βω −1 . Consequently, like T 5/2 , zrot like T , zvibr like sinh 2 ã ò ã ò ï Å 0 ï Å 0 z I tr z I2 tr d d ∆H =2 + ln zeI − + ln zrot + ln zvibr ln ln RT 2 dT NI dT NI2 (4.102) ò ï ò ï 5 Ed 1 ω βω 5 − + + + (4.103) coth =2 2 2 2T 2kB T 2T T 2kB T 2 that is, Å ã 3RT βω ω ∆H = + N Ed − coth 2 2 2

(4.104)

The obtained value of ∆H is close to that of the dissociation energy per mole : ∆H = 152 kJ/mole. The Kp and ∆H values, calculated from spectroscopic data, are very close to those deduced from the experiment quoted in the Appendix 4.1 (Kp 3.4 · 10−3 at 1000 K, ∆H = 158 kJ/mole). The discrepancies come in particular from the approximations done in the decomposition of the degrees of freedom of the iodine molecule : we assumed that the vibration and rotation were independent (rigid rotator model) ; moreover, the potential well describing vibration is certainly not harmonic. Obviously, for an experiment performed at constant temperature T and volume Ω, the equilibrium constant Kv could have been calculated in a similar way and the free energy F would have been used. The variation of Kv with temperature provides the variation ∆U of internal energy in the chemical reaction.

4. Conclusion Through these examples, which were developed in detail, the contribution of Statistical Physics in the prediction of chemical reactions was shown. As just

112

Appendix 4.2

seen now, if the Statistical Physics concepts refer to very fundamental notions, this discipline (sometimes at the price of lengthy calculations !) brings useful data to solve practical problems.

Chapter 5

Indistinguishability, the Pauli Principle (Quantum Mechanics) The preceding chapter presented properties of the ideal gas which are valid in the limit where the average value, at temperature T , of the de Broglie wavelength is small with respect to the average distance between the particles in motion. Another way to express the same concept is to consider that the wave functions of two particles have a very small probability of overlapping in such conditions. From now on, we will work in the other limit, in which collisions between particles of the same nature are very likely. In the present chapter the Quantum Mechanics properties applicable to this limit will be recalled. In Quantum Mechanics courses (see for example the one by J.-L. Basdevant and J. Dalibard) it is stated that, for indistinguishable particles, the Pauli principle must be added to the general postulates of Quantum Mechanics. In fact here we are only re-visiting this principle but ideas are better understood when they are tackled several times ! In chapter 6 the consequences in Statistical Physics will be drawn on the properties obtained in the present chapter 5. In the introduction (§ 5.1) we will list very diﬀerent physical phenomena, associated with the property of indistinguishability, which will be interpreted at the end of the chapter in the framework of Statistical Physics. To introduce the issue, in § 5.2 some properties of the quantum states of two, then several, identical particles, will be recalled. In § 5.3, the Pauli principle and 113

114

Chapter 5. Indistinguishability, the Pauli Principle

its particular expression, the Pauli exclusion principle, will be stated ; the theorem of spin-statistics connection will allow one to recognize which type of statistics is applicable according to the nature of the considered particles and to distinguish between fermions and bosons. In § 5.4, the special case of two identical particles of spin 1/2 will be analyzed. These results will be extended to an arbitrary number of particles in § 5.5 and the description of a N indistinguishable particle state through the occupation numbers of its various energy levels will be introduced : this is much more convenient than using the N -particle wave function. It is this approach that will be used in the following chapters of this book. Finally in § 5.6 we will return to the examples of the introduction and will interpret them in the framework of the Pauli principle.

5.1

Introduction

It is possible to distinguish ﬁxed particles of the same nature, by specifying their coordinates : this is the case of magnetic moments localized on sites of a paramagnetic crystal, of hydrogen molecules adsorbed on a solid catalyst. They are then distinguishable through their coordinates. D

θ (1)

(2)

a)

b)

c)

Fig. 5.1: Two identical particles, a) represented by the wave packets which describe them in their center-of-mass frame, b) suﬀer a collision. c) After the collision, the wave function is spread in a ring-shaped region the radius of which increases with time. When a particle is detected in D, it is impossible to know whether it was associated with wave packet (1) or (2) prior to the collision. On the other hand, two electrons in motion in a metal, two protons in a nucleus, or more generally all particles of the same chemical nature, identical, in motion and thus likely to collide, cannot be distinguished by their physical properties : these identical particles are said to be indistinguishable. Indeed, as explained in Quantum Mechanics courses, if these particles, the evolution of which is described through Quantum Mechanics, do collide, their wave

Introduction

115

functions overlap (ﬁg. 5.1). Then, after the collision it is impossible to know whether (1 ) originated from (1) or from (2), since the notion of trajectory does not exist in Quantum Mechanics (Fig. 5.2). D

D (1 )

(2 )

(1)

(1) (2)

(2) (2 )

(1 ) a)

b)

Fig. 5.2: Representation in term of “trajectory” of the event described on Fig. 5.1. The two particles being identical, one cannot distinguish between a) and b). The quantum treatment of such systems refers to the Pauli principle. This principle allows the interpretation of a great variety of chemical and physical phenomena, among which are 1. the periodic classiﬁcation of the elements ; 2. the chemical binding ; 3. the ferromagnetism of some metals, like iron or cobalt, that is, their ability to evidence a macroscopic magnetization even in the absence of an external magnetic ﬁeld ; 4. the superconductivity of some solids, the electrical resistance of which strictly vanishes below a temperature called “critical temperature” ; 5. the special behavior of the isotope 42 He, which becomes superﬂuid for temperatures lower than a critical temperature Tc , equal to 2.17 K under the atmospheric pressure : this means that no viscosity is restraining its motion ; the isotope 32 He does not show such a property ; 6. the stimulated light emission and the laser eﬀect.

116

5.2

5.2.1

Chapter 5. Indistinguishability, the Pauli Principle

States of Two Indistinguishable Particles (Quantum Mechanics) General Case

When N particles are indistinguishable, the physics of the system and the N particle hamiltonian are unchanged by permutation of some of these particles. Consequently, the hamiltonian commutes with any permutation operator between these particles, so that they have a common basis of eigenvectors. It can be shown that any permutation of N particles can be decomposed into a product of exchanges of two particles. Thus it is helpful to begin by recalling the properties of the operator which exchanges the particles of a system consisting in particle 1 at r1 and particle 2 at r2 : Pˆ12 ψ(r1 , r2 ) = ψ(r2 , r1 )

(5.1)

The operator Pˆ12 must satisfy : (Pˆ12 )2 = 1

(5.2)

since after two exchanges one returns to the initial state. This implies that the eigenvalue eiα of Pˆ12 must fulﬁll the condition : e2iα = 1, i.e., eiα = ±1

(5.3)

eiα = +1, ψ(r2 , r1 ) = ψ(r1 , r2 )

(5.4)

i) If

the eigenstate is symmetrical in the exchange (or transposition) of particles 1 and 2. ii) If eiα = −1, ψ(r2 , r1 ) = −ψ(r1 , r2 )

(5.5)

the wave function is antisymmetrical in the transposition of the two particles.

5.2.2

Independent Particles

If the indistinguishable particles are independent, the total hamiltonian for all the particles is written as a sum of one-particle hamiltonians : ˆN = H

N i=1

ˆi h

(5.6)

States of two Indistinguishable Particles (Quantum Mechanics)

117

The eigenstates of each one-particle hamiltonian satisfy : ˆ i χn (ri ) = εn χn (ri ) (i = 1, 2) h

(5.7)

The index n corresponds to the considered quantum state, for example for the electron of a hydrogen atom n = 1s, 2s, 2p . . . ; the particle i is located at ri . The one-particle eigenstate at the lowest energy (fundamental state) satisﬁes ˆ 1 χ1 (r1 ) = ε1 χ1 (r1 ) h

(5.8)

The energy ε2 of the one-particle ﬁrst excited state is such that ˆ 1 χ2 (r1 ) = ε2 χ2 (r1 ) h

(5.9)

Since the particles are assumed to be independent, the eigenstates of the ˆ 12 can be easily expressed with respect to the onetwo-particle hamiltonian H particle solutions : ˆ 12 ψnn (r1 , r2 ) = εnn ψnn (r1 , r2 ) H

(5.10)

εnn = εn + εn and ψnn (r1 , r2 ) = χn (rn )χn (r2 )

(5.11)

with

Indeed remember that ˆ1 + h ˆ 2 )χn (r1 )χn (r2 ) = εn χn (r1 )χn (r2 ) + χn (r1 )εn χn (r2 ) (h = (εn + εn )χn (r1 )χn (r2 )

(5.12)

so that the two-particle energy is the sum of the one-particle energies and the two-particle wave function is the product of the one-particle wave functions. Any linear combination of wave functions corresponding to the same energy ˆ 12 . εn + εn is also a solution of H Here one is looking for eigenfunctions with a symmetry determined by the indistinguishability properties. In the two-particle fundamental state, each individual particle is in its fundamental state (ﬁg. 5.3), this two-particle state has the energy 2ε1 , it is nondegenerate. Its wave function satisﬁes ˆ 12 ψ11sym (r1 , r2 ) = 2ε1 ψ11sym (r1 , r2 ) H

(5.13)

It is symmetrical and can be written ψ11sym (r1 , r2 ) = χ1 (r1 )χ1 (r2 ) .

(5.14)

118

Chapter 5. Indistinguishability, the Pauli Principle

ε2

ε2

ε1

ε1

fundamental state

ﬁrst excited state

Fig. 5.3 : Occupation of two levels by two indistinguishable particles. In the two-particle ﬁrst excited state, the total energy is ε1 + ε2 . This state is doubly degenerate, as two distinct independent states can be built (particle 1 in ε1 , particle 2 in ε2 or the opposite). But since the particles are indistinguishable, one cannot tell which of the two particles is in the fundamental state. ˆ 12 and Pˆ12 . Thus one looks for solutions which are common eigenstates of H The acceptable solutions are : for the symmetrical ﬁrst excited state 1 ψ12sym (r1 , r2 ) = √ (χ1 (r1 )χ2 (r2 ) + χ2 (r1 )χ1 (r2 )) 2

(5.15)

for the antisymmetrical ﬁrst excited state 1 ψ12asym (r1 , r2 ) = √ (χ1 (r1 )χ2 (r2 ) − χ2 (r1 )χ1 (r2 )) 2

(5.16)

These wave functions (5.14) to (5.16) are determined, to a phase factor. Now consider the case of N particles. ˆ N , is then equal The total energy of the system, which is the eigenvalue of H to εN =

N

εni

(5.17)

i=1

The eigenfunction is a product of one-particle eigenfunctions, N

ψN (r1 , . . . , rN ) = Π χni (ri ) i=1

(5.18)

Any permutation of N particles can be decomposed into a product of exchanges of particles by pair, it thus admits the eigenvalues +1 or −1 : Pˆ ψ(r1 , r2 , . . . ri , . . . rN ) = σψ(r1 , r2 , . . . ri , . . . , rN )

(5.19)

Pauli Principle ; Spin-Statistics Connection

119

with σ = ±1. σ = +1 is associated with an “even” permutation σ = −1 is associated with an “odd” permutation Although the decomposition of the permutation into a product of exchanges is not unique, the value of σ is well determined. Among the eigenfunctions of the permutation operators, two types will be distinguished : either the wave function is completely symmetrical, i.e., it is unchanged by any permutation operator ; or it is a completely antisymmetrical eigenfunction which satisﬁes (5.19) and changes its sign for an odd permutation, but is not changed for an even permutation. The eigenfunctions, common ˆ N of the N indistinto the permutation operators and to the hamiltonian H guishable particles, will necessarily be of one of these two types.

5.3

5.3.1

Pauli Principle ; Spin-Statistics Connection Pauli Principle ; Pauli Exclusion Principle

The Pauli (1900-1958) principle is a postulate which is stated as follows : The wave function of a set of N indistinguishable particles obeys either of the two properties below : – either it is completely symmetrical, i.e., its sign does not change by permutation of two arbitrary particles : then these particles are called bosons ; – or it is completely antisymmetrical, and thus its sign is modiﬁed by permutation of two particles : these particles are fermions. The property veriﬁed by the wave function only depends on the type of the considered particles, it is independent of their number N and of the particular state Ψ they occupy. This principle is veriﬁed by all chemical types of particles. In particular, every time a new type of elementary particle was discovered, these new particles indeed satisﬁed the Pauli principle. A special case of the Pauli principle is the Pauli exclusion principle : Two fermions cannot be in the same quantum state. Indeed, if they were in the

120

Chapter 5. Indistinguishability, the Pauli Principle

same one-particle state Ψ, the antisymmetrical expression of the wave function would be : 1 ψ(1, 2) = √ [ψ(1)ψ(2) − ψ(2)ψ(1)] = 0 2

(5.20)

One now needs to determine which particles, called “fermions”, satisfy the Fermi-Dirac statistics, and which particles, the “bosons”, follow the BoseEinstein statistics. There is no intuitive justiﬁcation to the following theorem, which provides the answer to this question.

5.3.2

Theorem of Spin-Statistics Connection

The particles with an integral or zero spin (S/ = 0, 1, 2 . . . ) are bosons. The wave function for N bosons is symmetrical in the transposition of two of them. The particles with an half-integral spin (S/ = 1/2, 3/2, 5/2 . . . ) are fermions. The N -fermions wave function is antisymmetrical in the transposition of two of them. Among the elementary particles, electrons, protons, neutrons, neutrinos, muons have a spin equal to 1/2 ; the spin of photons is 1, the spin of mesons is 0 or 1. When these particles are combined into one or several nuclei, the total angular momentum takes into account all the orbital momenta and the spin intrinsic momenta. In Quantum Mechanics courses it is shown that the addition of two angular momenta J1 and J2 gives a parameter which has the properties of a new angular momentum, with a quantum number J varying between | J1 −J2 | and | J1 + J2 | by unity steps. If two integral or half-integral momenta are added, only integral values of J are obtained ; if an integral J1 is added to a half-integral J2 , the sum J is half-integral. It is the quantum number J which is identiﬁed with the spin of the ensemble. More generally if a composite particle contains an even number of elementary particles of half-integral spin, its total angular moment (or its total spin) is integral, this is a boson ; if it contains an odd number, its total angular moment is half-integral, and the composite particle is a fermion. For example, the atom 42 He has two electrons, two protons and two neutrons, i.e., six particles of half-integral spin : it is a boson. On the other hand, the atom 32 He is composed of two electrons, two protons but a single neutron, thus it is a fermion. The statistical properties of these two isotopes are very diﬀerent (see the Introduction of this chapter, example 5), whereas the che-

Case of Two Particles of Spin 1/2

121

mical properties, determined by the number of electrons that can participate in chemical bonds, are the same.

5.4

Case of Two Particles of Spin 1/2

Here the properties of the wave functions of two indistinguishable particles of spin 1/2 are recalled, this is in particular the case for two electrons. From the theorem of spin-statistics connection, this wave function must be antisymmetrical with respect to the transposition of these two particles. Now, in addition to their spin degree of freedom, these two particles also have an orbital degree of freedom concerning the space coordinates ; from Quantum Mechanics courses, the wave function for a single particle of this type can be decomposed into tensor products concerning the space and spin variables. The most general form of decomposition is : ψ(r1 , r2 ; σ1 , σ2 ) = ψ++ (r1 , r2 )| + + + ψ+− (r1 , r2 )| + − + ψ−+ (r1 , r2 )| − + + ψ−− (r1 , r2 )| − −

5.4.1

(5.21)

Triplet and Singlet Spin States

Let us ﬁrst assume, for simpliﬁcation, that this decomposition is reduced to a single term, i.e., ψ(r1 , r2 ; σ1 , σ2 ) = ψ(r1 , r2 ) ⊗ |σ1 , σ2 , with σ1 , σ2 = ±1/2

(5.22)

Since we are dealing with fermions, the two-particle wave function changes sign when particles 1 and 2 are exchanged. This sign change can occur in two diﬀerent ways : – either it is the space part ψ(r1 , r2 ) of the wave function which is antisymmetrical in this exchange, the spin function remaining unchanged ; – or it is |σ1 , σ2 which changes its sign in this exchange, the space function remaining unchanged. Consider the spin states. Each particle has two possible spin states, noted | ↑ or | ↓. There are thus four basis states for the set of two spins, which can be chosen as | ↑↑, | ↑↓, | ↓↑, | ↓↓ (the order of notation is : particle 1, then particle 2).

(5.23)

122

Chapter 5. Indistinguishability, the Pauli Principle

From these states one deduces four new independent states, classiﬁed according to their symmetry, that is, their behavior when exchanging the spins of the two particles. There are three symmetrical states, unchanged when the role of particles 1 and 2 is exchanged : ⎧ ⎪ | ↑↑ = |11 ⎪ ⎪ ⎪ ⎨ 1

√ | ↑↓ + | ↓↑ = |10 ⎪ ⎪ 2 ⎪ ⎪ ⎩ | ↓↓ = |1 − 1

(5.24)

They constitute the triplet state of total spin S = 1, with projections on the quantization axis Oz described by the respective quantum numbers Sz = +1, 0, −1 : the eigenvalue of the squared length of this spin is equal to S(S + 1)2 , i.e., 22 when the eigenvalue of the projection along Oz is Sz , i.e., , 0 or −. There remains an antisymmetrical state, which changes its sign in the exchange of the two spins, this is the singlet state 1

√ | ↑↓ − | ↓↑ = |00 2

(5.25)

in which S = Sz = 0. Even when the two-particle hamiltonian does not explicitly depend on spin, orbital wave functions with diﬀerent symmetries and possibly diﬀerent energy eigenvalues are associated with the triplet and singlet states : for example, in the special case treated in § 5.2.2 the fundamental state, with a symmetrical orbital wave function [see Eq. (5.14)], must be a spin singlet state, whereas in the ﬁrst excited state one can have a spin triplet. This energy splitting between the symmetrical and antisymmetrical orbital solutions is called the exchange energy or exchange interaction. The exchange interaction is responsible of ferromagnetism, as understood by Werner Heisenberg (1901-1976) as early as 1926 : indeed the dipolar interaction between two magnetic moments a fraction of a nm apart, each of the order of the Bohr magneton, is much too weak to allow the existence of an ordered phase of magnetization at room temperature. On the contrary, owing to the Pauli principle, in ferromagnetic systems the energy splittings between states with diﬀerent spins correspond to separations between diﬀerent orbital levels. These are of a fraction of an electron-volt, which is equivalent to temperatures of the order of 1000 K, since these states are the ones responsible for valence bindings.

Case of Two Particles of Spin 1/2

123

In a ferromagnetic solid, magnetism can originate from mobile electrons in metals like in Fe, Ni, Co, or from ions in insulating solids, as in the magnetite iron oxide Fe3 O4 , and the spin can diﬀer from 1/2. However this interaction between the spins, which relies on the Pauli principle, is responsible for the energy splitting between diﬀerent orbital states : one state in which the total magnetic moment is diﬀerent from zero, corresponding to the lower temperature state, and the other one in which it is zero.

5.4.2

General Properties of the Wave Function of Two Spin 1/2 Particles

It has just been shown that, because of the Pauli principle, space wave functions are associated with the triplet states, which are changed into their opposite in the exchange of the two particles (antisymmetrical wave function). To the singlet state is associated a symmetrical wave function. The most general wave function for two spin 1/2 particles will be written, for a decomposition onto the triplet and singlet states : A A A S |11 + ψ10 |10 + ψ1−1 |1 − 1 + ψ00 |00 ψ11 A The wave functions of the type ψ11 are spatial functions of r1 and r2 , the upper index recalls the symmetry of each component.

The same wave function can be expressed on the spin basis |σ1 , σ2 , accounting for expressions (5.24) and (5.25) of the triplet and singlet wave functions. One thus obtains : A ψ(r, r2 ; σ1 , σ2 ) = ψ11 | + + + A + ψ1−1 | − −

A ψA − ψS + ψS ψ10 √ 00 | + − + 10√ 00 | − + 2 2

(5.26)

It thus appears, by comparing to the general expression (5.21), that the Pauli principle dictates symmetry conditions on the spin 1/2 components : • ψ++ (r1 , r2 ) and ψ−− (r1 , r2 ) must be antisymmetrical in the exchange of particles 1 and 2 • ψ+− (r1 , r2 ) and ψ−+ (r1 , r2 ) do not have a speciﬁc symmetry, but their sum is antisymmetrical and their diﬀerence symmetrical.

124

5.5

Chapter 5. Indistinguishability, the Pauli Principle

Special Case of N Independent Particles ; Occupation Numbers of the States

In the general case of an N -identical-particle system, the Pauli principle dictates a form for the N -particle wave function and occupation conditions for the energy levels which are very diﬀerent in the case of fermions or bosons. In this part of the course, we are only interested by independent particles : this situation is simpler and already allows a large number of physical problems to be solved.

5.5.1

Wave Function

Fermions : We have seen that two fermions cannot be in the same quantum state and that their wave function must be antisymmetrical in the transposition of the two particles. We consider independent particles, the N -particle wave function of which takes the form of a determinant called “Slater determinant,” which indeed ensures the change of sign by transposition and the nullity if two particles are in the same state : χ1 (r1 ) χ2 (r1 ) χN (r1 ) .. ... ψ(r1 , . . . , rN ) ∝ ... . .. χ1 (rN ) . χN (rN )

(5.27)

If a state is already occupied by a fermion, another one cannot be added. A consequence is that at zero temperature all the particles cannot be in the minimum energy state : the diﬀerent accessible levels must be ﬁlled, beginning with the lowest one, each time putting one particle per level and thus each time moving to the next higher energy level. The process stops when the N th and last particle is placed in the last occupied level : in Chemistry this process is called the “Aufbau principle” or construction principle ; in solids, the last occupied state at zero temperature is called the Fermi level and is usually noted εF (see chapters 7 and 8). Bosons : On the other hand, the wave function for N bosons is symmetrical ; in particular it is possible to have all the particles on the same level of minimum energy at zero temperature.

Special Case of N Independent Particles ; Occupation Numbers

5.5.2

125

Occupation Numbers

The N -fermion (or N -boson) wave functions are complicated to write because, as soon as a particle is put in a state of given energy, this induces conditions on the available states for the other particles ; then the antisymmetrization of the total wave function has to be realized for the fermions (or the symmetrization for the bosons). Since these particles are indistinguishable, the only interesting datum in Statistical Physics is the number of particles in a deﬁned quantum state. Instead of writing the N -particle wave function, which will be of no use in what follows, we will described the system as a whole by the occupation numbers of its diﬀerent states, this is the so-called “Fock basis” : these one-particle states correspond to distinct εk , which can have the same energy but diﬀerent indexes if the state is degenerate (for example two states of the same energy and diﬀerent spins will have diﬀerent indexes k and k ) ; the states are ordered by increasing energies (Fig. 5.4). εk

nk

ε3

n3

ε2

n2

ε1

n1

Fig. 5.4 : Deﬁnition of the occupation numbers of the one-particle states. The notation |n1 ,2 , . . . nk , . . . will express the N -particle state in which n1 particles occupy the state of energy ε1 , n2 that of energy ε2 , . . . nk the state of energy εk . One must satisfy n1 + n2 + . . . + nk + . . . = N

(5.28)

the total number of particles in the system. Under these conditions, the total energy of the N -particle system is given by n1 ε1 + n2 ε2 + . . . + nk εk + . . . = EN

(5.29)

From the Pauli principle • for fermions : nk = 0 or 1

(5.30)

• for bosons

(5.31)

: nk = 0, 1, 2 . . . ∞

126

Chapter 5. Indistinguishability, the Pauli Principle

This description will be used in the remainder of the course. The whole physical information is included in the Fock notation. Indeed one cannot tell which particle is in which state, this would be meaningless since the particles are indistinguishable. The only thing that can be speciﬁed is how many particles are in each one-particle state.

5.6

Return to the Introduction Examples

Now we are able to propose a preliminary interpretation of the physical phenomena described in the introduction, § 5.1, some of which will be analyzed in more detail in the following chapters :

5.6.1

Fermions Properties

1. In the periodic classiﬁcation of the elements, the atomic levels are ﬁlled with electrons which are fermions. In a given orbital level only two electrons of distinct spin states can be placed, i.e., according to the Pauli principle they are in the singlet spin state. 2. When a chemical bond is created, the electrons coming from the two atoms of the bond go into the atomic levels originating from the coupling between the levels of the separated atoms. The resulting bonding and antibonding states must also be populated by electrons in diﬀerent spin states, thus in the singlet spin state. 3. The ferromagnetism arises, as already mentioned, from the energy difference between the singlet and the triplet spin states for a pair of electrons. If the fundamental state presents a macroscopic magnetization, it must be made from microscopic states with a nonzero spin. In fact, in a ferromagnetic metal such as iron or nickel, the spins of two iron nuclei couple through conduction electrons which travel from nucleus to nucleus.

5.6.2

Bosons Properties

4. In superconductivity the electrons of a solid couple by pairs due to their interaction with the solid vibrations. These pairs, called “Cooper pairs,” are bosons which gather into the same fundamental state at low temperature, thus producing the phenomenon of superconductivity. 5. In the same way the 4 He atoms are bosons, which can occupy the same quantum state below a critical temperature (see chapter 9). Their ma-

Return to the Introduction Examples

127

croscopic behavior expresses this collective property, whence the absence of viscosity. 6. Photons are also bosons and the stimulated emission, a speciﬁc property of bosons in nonconserved number, will be studied in § 9.2.4.

Summary of Chapter 5 In Quantum Mechanics, the hamiltonian describing the states of identical particles is invariant under a permutation of these particles, the states before and after permutation are physically indistinguishable. This means that the hamiltonian commutes with any permutation operator : one ﬁrst looks for the permutations eigenstates, which are also eigenstates of the N -particle hamiltonian. In the case of two indistinguishable particles the symmetrical and antisymmetrical eigenstates have been described. These states have a simple expression in the case of independent particles. Any permutation of N particles can be decomposed into a product of exchanges (or transpositions) of particles by pair : a permutation is thus either even or odd. The Pauli principle postulates that the wave functions for N identical particles can only be, according to the particle’s nature, either completely symmetrical (bosons) or completely antisymmetrical (fermions) in a permutation of the particles. In particular, because their wave function is antisymmetrical, two fermions cannot occupy the same quantum state (“Pauli exclusion principle”). The theorem of spin-statistics connection speciﬁes that the particles of halfintegral total spin, the fermions, have an antisymmetrical wave function ; the particles of integral or zero total spin, the bosons, have a completely symmetrical wave function. In the case of two spin-1/2 particles, the triplet spin state, symmetrical, and the singlet spin state, antisymmetrical, have been described. The spin component of the wave function has to be combined with the orbital part, the global wave function being antisymmetrical to satisfy the Pauli principle. In Statistical Physics, to account for the Pauli principle, the state of N particles will be described using the occupation numbers of the various energy 129

130

Summary of Chapter 5

levels, rather than by its wave function. A given state can be occupied at most by one fermion, whereas the number of bosons that can be located in a given state is arbitrary.

Chapter 6

General Properties of the Quantum Statistics In chapter 5 we learnt the consequences, in Quantum Mechanics, of the Pauli principle which applies to indistinguishable particles : they will now always be taken as independent, that is, without any mutual interaction. We saw that it is very diﬃcult to express the conditions stated by the Pauli principle on the N -particle wave function ; on the other hand, a description through occupation numbers |n1 , n2 , . . . , nk , . . . of the diﬀerent quantum states ε1 , ε2 , . . . εk . . . accessible to these indistinguishable particles is much more convenient. In fact we assume that the states are nondegenerate and if there are several states with the same energy, we label them with diﬀerent indexes. Besides, it was stated that there exists two types of particles, the fermions with a half-integral spin, which cannot be more than one in any given one-particle state εk , whereas the bosons, with an integral or zero spin, can be in arbitrary number in any one single-particle quantum state. In the present chapter we return to Statistical Physics (except in § 6.4, which is on Quantum Mechanics) : after having presented the study technique applicable to systems referring to Quantum Statistics, we determine the average number of particles, at a given temperature, on an energy state of the considered system : this number will be very diﬀerent for fermions and bosons. This chapter presents general methods and techniques to analyze, using Statistical Physics, the properties of systems of indistinguishable particles following the Quantum Statistics of either Fermi-Dirac or Bose-Einstein, while the following chapters will study these statistics more in detail on physical examples.

131

132

Chapter 6. General Properties of the Quantum Statistics

In § 6.1, we will show how the statistical properties of indistinguishable independent particles are more conveniently described in the grand canonical ensemble. In § 6.2 we will explain how the grand canonical partition function ZG is factorized into terms, each one corresponding to a one-particle quantum state. In § 6.3 the Fermi-Dirac and the Bose-Einstein distributions will be deduced, which give the average number of particles of the system occupying a quantum state of given energy at temperature T , and the associated thermodynamical potential, the grand potential, will be expressed. The § 6.4 will be a (long) Quantum Mechanics parenthesis, it will be shown there that a macroscopic system, with extremely close energy levels, is conveniently described by a density of states, that will be calculated for a free particle. In § 6.5, this density of states will be used to obtain the average value, at ﬁxed temperature, of physical parameters in the case of indistinguishable particles. Finally, in § 6.6, the criterion used in § 4.3 to deﬁne the domain of Classical Statistics will be justiﬁed and it will be shown that the Maxwell-Boltzmann Classical Statistics is the common limit of both Quantum Statistics when the density decreases or the temperature increases.

6.1

Use of the Grand Canonical Ensemble

In chapter 5 we learnt the properties requested for the wave function of N indistinguishable particles, when these particles are fermions or when they are bosons. (From the present chapter until the end of the book, the study will be limited to indistinguishable and independent particles.) Now we look for the statistical description of such a system in thermal equilibrium at temperature T and ﬁrst show, on the example of two particles, that the use of the canonical ensemble is not convenient.

6.1.1

Two Indistinguishable Particles in Thermal Equilibrium : Statistical Description in the Canonical Ensemble

Consider a system of two indistinguishable particles. The one-particle energy levels are noted ε1 , . . . , εk , . . . , and are assumed to be nondegenerate (a degenerate level is described by several states at the same energy but with diﬀerent indexes). The canonical partition function for a single particle at temperature T is written : 1 (6.1) e−βεk with β = Z1 = kB T k

Use of the Grand Canonical Ensemble

133

Now consider the two particles. For a given microscopic conﬁguration, the total energy is εi + εj . After having recalled the results for two distinguishable particles, we will have to separately consider the case of fermions and that of bosons. We have seen in § 2.4.5 that for distinguishable particles Z2disc = (Z1 )2

(6.2)

since there is no limitation to the levels occupation by the second particle. Fermions cannot be more than one to occupy the same state. Consequently, as soon as the level of energy εi is occupied by the ﬁrst particle, the second one cannot be on it. Thus 1 −β(εi +εj ) −β(εi +εj ) Z2fermions = e = e (6.3) 2 i<j i=j

Let us compare Z2fermions to (Z1 )2 : 2 −βεi −βεj (Z1 ) = e e e−β2εi + 2 e−β(εi +εj ) = i

(Z1 )2 =

j

i

(6.4)

i<j

e−β2εi + 2Z2fermions

(6.5)

i

In the case of bosons, in addition to the conﬁgurations possible for the fermions, the two particles can also lie on the same level : e−β(εi +εj ) (6.6) Z2bosons = i≤j 2

Let us relate (Z1 ) to Z2bosons : (Z1 )2 = Z2bosons +

1 −β(εi +εj ) e 2

(6.7)

i=j

(Z1 )2 = Z2bosons + Z2fermions

(6.8)

It was easy to determine the levels available for the particles in this example because they were only two. For three particles one must then consider which restrictions appear for the third one because of the levels occupied by the ﬁrst two particles. For N particles one should again proceed step by step, but this method is no longer realistic !

6.1.2

Description in the Grand Canonical Ensemble

To avoid complicated conditions on the occupation numbers nk one will work in the grand canonical ensemble, assuming in a ﬁrst step that the total number

134

Chapter 6. General Properties of the Quantum Statistics

of particles of the system is arbitrary. Thus one will calculate ZG and the grand potential A ; then the value of the Lagrange parameter α = βµ (or that of the chemical potential µ) will be ﬁxed in such a way that the average number of particles N , deduced from the statistical calculation, coincides with the real number of particles in the system. We know (§ 2.5.4) that the √ relative ﬂuctuation on N introduced by this procedure is of the order of 1/ N , that is, of the order of a few 10−12 for N of the order of the Avogadro number.

6.2

6.2.1

Factorization of the Grand Partition Function Fermions and Bosons

By deﬁnition ∞

ZG (α, β) =

∞

eαN ZN (β) =

N =0

eβµN ZN (β)

(6.9)

N =0

where N is one of the numbers of particles and where α and the chemical potential are related by α = βµ. The canonical partition function ZN (β) corresponding to N particles is given by

ZN (β) =

e−βEn

(6.10)

n

Let us introduce the occupation numbers nk of the quantum states εk :

ZG =

∞

βµ

e

Ñ

nk

k

N =0

−β

e

é

nk εk

k

(6.11)

nk =N

with

nk = N

(6.12)

k

and the energy for this conﬁguration of the N particles being given by k

nk εk = EN

(6.13)

Factorization of the Grand Partition Function

135

As the sum is performed over all the total numbers N and all the particle distributions in the system, ZG is also written βµ nk −β nk εk k ZG = e k (6.14) {|n1 ...nk >}

=

β

e

nk (µ−εk )

k

(6.15)

{|n1 ...nk >}

In the latter expression, one separates the contribution of the state ε1 : this is the term eβ(µ−ε1 )n1 (6.16) n1

which multiplies the sum concerning all the other states, with n1 taking all possible values. One then proceeds state after state, so that ZG is written as a product of factors : ∞ β(µ−εk )nk ZG = e (6.17) k=1

nk

Each factor concerns a single state εk and its value depends on the number of particles which can be located in this state, i.e., of the nature, fermions or bosons, of the considered particles. Note that, since ZG is a product of factors, ln ZG is a sum of terms, each concerning a one-particle state of the type εk .

6.2.2

Fermions

In this case, the quantum state εk is occupied by 0 or 1 particle : there are only two terms in nk , the contribution of the state εk in ZG is the factor 1 + exp β(µ − εk ), whence ZG fermions =

∞

(1 + eβ(µ−εk ) )

(6.18)

k=1

6.2.3

Bosons

The state εk can now be occupied by 0, 1, 2,... particles, so that nk varies from zero to inﬁnity. The contribution in ZG of the state εk is thus the factor ∞

eβ(µ−εk )nk =

nk =0

1 1 − eβ(µ−εk )

(6.19)

136

Chapter 6. General Properties of the Quantum Statistics

This geometrical series can be summed only if exp(α − βεk ) is smaller than unity, a condition to be fulﬁlled by all the states εk . One will thus have to verify that µ < ε1 , where ε1 is the one-particle quantum state with the lowest energy, i.e., the fundamental state. This condition is also expressed through α = βµ as α < βε1 . Then the grand partition function takes the form : ZG bosons =

∞ k=1

6.2.4

1 1−

eβ(µ−εk )

(6.20)

Chemical Potential and Number of Particles

The grand canonical partition function ZG is associated with the grand potential A such that A = −kB T ln ZG

(6.21)

As speciﬁed in § 3.5.3, the partial derivatives of the grand partition function provide the entropy, the pressure and the average number of particles N : dA = −SdT − P dΩ − N dµ

(6.22)

In particular the constraint on the total number of particles of the system is expressed from : Å ã ∂A N =− (6.23) ∂µ T,Ω

6.3

Average Occupation Number ; Grand Potential

From the grand partition function ZG one deduces average values at temperature T of physical parameters of the whole system, but also of the state εk . In particular the average value of the occupation number nk of the state of energy εk is obtained through nk =

nk eβ(µ−εk )nk 1 ∂ ln ZG β(µ−ε )n =− k k β ∂εk e

nk

nk

(6.24)

Average Occupation Number ; Grand Potential

137

For fermions, the average occupation number, the so-called “Fermi-Dirac distribution,” is given by nk F D =

1 1 = βε −α e k +1 eβ(εk −µ) + 1

with α = βµ

(6.25)

For bosons, this average number is “the Bose-Einstein” distribution which takes the form nk BE =

1 1 = βε −α e k −1 eβ(εk −µ) − 1

(6.26)

Note : Remember the sign diﬀerence in the denominator between fermions and bosons ! Its essential physical consequences are developed in the next chapters. The average occupation numbers are related to factors in the grand partition function ZG : indeed one will verify that 1 = 1 + eβ(µ−εk ) 1 − nk F D 1 1 + nk BE = 1 − eβ(µ−εk )

(6.27) (6.28)

which are the respective contributions of the level εk to ZG , in the cases of fermions or bosons. These occupation numbers yield the average values, at given temperature, thus at ﬁxed β, of physical parameters. Thus the average number of particles is related to α and β, which appear in f (ε), through N =

nk

(6.29)

k

The total energy of the system at temperature T is obtained from U=

εk nk

(6.30)

k

The grand potential A is expressed, like the ZG factors, versus the occupation numbers of the states of energy εk : A = +kB T

ln(1 − nk ) fermions

(6.31)

ln(1 + nk ) bosons

(6.32)

k

A = −kB T

k

138

6.4

Chapter 6. General Properties of the Quantum Statistics

Free Particle in a Box ; Density of States (Quantum Mechanics)

The diﬀerent physical parameters, like the average number of particles N , the internal energy U , the grand potential A, were expressed above as sums of contributions arising from the discrete one-particle levels. Now we will ﬁrst recall what these states are for a free particle conﬁned within a volume of macroscopic characteristic dimensions, i.e., of the order of a mm or a cm (§ 6.4.1) : such dimensions are extremely large with respect to atomic distances, so that the characteristic energy splittings are very small (§ 6.4.1a). The boundary conditions (§ 6.4.1b) of the wave function imply quantization conditions, that we are going to express in two diﬀerent ways. However, for both quantization conditions the same energy density of states D(ε) is deﬁned (§ 6.4.2), such that the number of allowed states between energies ε and ε + dε is equal to D(ε) dε.

6.4.1

Quantum States of a Free Particle in a Box

a) Eigenstates By deﬁnition a free particle feels no potential energy (except the one expressing the possible conﬁnement) We ﬁrst consider that the free particle can move through the entire space. Its hamiltonian is given by pˆ 2 ˆ h= 2m

(6.33)

We are looking for its stationary states (see a course on Quantum Mechanics) : the eigenstate |ψ and the corresponding energy ε satisfy pˆ 2 |ψ = ε|ψ 2m

(6.34)

i.e., 2 ∆ψ(r) = εψ(r) 2m The time-dependent wave function is then deduced : −

Ψ(r, t) = ψ(r) exp(−iεt/)

(6.35)

(6.36)

It is known that the three space variables separate in eq. (6.35) and that it is enough to solve the one-dimension Schroedinger equation −

2 d2 ψ(x) = εx ψ(x) 2m dx2

(6.37)

Free Particle in a Box ; Density of States

139

which admits two equivalent types of solutions : i) either 1 ψ(x) = √ exp ikx x Lx with kx > 0, < 0 or zero. The time-dependent wave function Å ã 1 εx t Ψ(x, t) = √ exp i kx x − Lx

(6.38)

(6.39)

2 kx2 = ωx is a progressive wave : when the time t increases, with εx = 2m a state of given phase propagates toward increasing x’s for kx > 0, toward decreasing x’s for kx < 0 ; the space wave function is a constant for kx = 0. ii) or ψ(x) = A cos kx x + B sin kx x

(6.40)

where A and B are constants. The time-dependent wave function ã Å εx t Ψ(x, t) = (A sin kx x + B cos kx x) exp −i

(6.41)

corresponds to the same energy εx or the same pulsation ωx as in (6.39). The solution of progressive-wave type (6.39) is found again if one chooses √ A = 1/ Lx , B = iA. For real coeﬃcients A and B, when changing the space origin in x there is indeed separation of the space and time variables in (6.41). Then a state of given space phase does not propagate in time, such a wave is stationary. At three dimensions the kinetic energy terms along the three coordinates add up in the hamiltonian ; the energy is equal to ε = εx + εy + εz

(6.42)

and the wave function is the product of three wave functions, each concerning a single coordinate. Now we are going to express, through boundary conditions on the wave function, that the particle is conﬁned within the volume Ω. This will yield quantization conditions on the wave vector and the energy.

140

Chapter 6. General Properties of the Quantum Statistics

b) Boundary Conditions The volume Ω containing the particles has a priori an arbitrary shape. However, it can be understood that each particle in motion is most of the time at a large distance from the container walls, so that the volume properties are not sensitive to the surface properties as soon as the volume is large enough. It is shown, and we will admit it, that the Statistical Physics properties of a macroscopic system are indeed independent of the shape of the volume Ω. For convenience we will now assume that Ω is a box (Lx , Ly , Lz ), with macroscopic dimensions of the order of a mm or a cm. We consider a one-dimension problem, on a segment of length Lx . The presence probability of the particle is zero outside the interval [O, Lx ], in the region where the particle cannot be found. There are two ways to express this condition : i) Stationary boundary conditions : To indicate that the particle cannot leave the interval, one assumes that in x = 0 and x = Lx potential barriers are present, which are inﬁnitely high and thus impossible to overcome. Thus the wave function vanishes outside the interval [O, Lx ] and also at the extremities of the segment, to ensure its continuity at these points ; since it vanishes at x = 0, it must be of the form ψ(x) = A sin kx x (Fig. 6.1). The cancellation at x = Lx yields kx Lx = nx π

(6.43)

where nx is an integer, so that π Lx π The wave vector is an integer multiple of , it is quantized. Lx kx = nx

For such values of kx , the time-dependent wave function is given by ã Å εx t Ψ(x, t) = A sin kx x exp −i

(6.44)

(6.45)

One understands that the physically distinct allowed values are restricted to nx > 0 : indeed taking nx < 0 is just changing the sign of the wave function, which does not change the physics of the problem. The value nx = 0 is to be rejected since a wave function has to be normalized to unity and thus cannot cancel everywhere. In the same way ky and kz are quantized by the conditions of cancellation of the wave function on the surface : ﬁnally the allowed vectors k for the particle conﬁned within the volume Ω are of the form ã Å k = nx π , ny π , nz π (6.46) Lx Ly Lz

Free Particle in a Box ; Density of States

141

ψ(x)

x O

Lx

Fig. 6.1: The free-particle wave function is zero outside the interval ]O, Lx [. the three integers nx , ny , nz being strictly positive. In the space of the wave vectors (Fig. 6.2), the extremities ã are on a rectangular lattice Å of these vectors π π π k . The unit cell built on these deﬁned by the three vectors i, j, Lx Ly Lz π3 π3 . = vectors is Lx Ly Lz Ω kz

L

x

π

π Ly π Lz

ky

kx Fig. 6.2: Quantization of the k-space for boundary conditions of the stationary-wave type. Only the trihedral for which the three k-components are positive is to be considered. ii) Periodical boundary [Born-Von Kármán (B-VK)] conditions : Here again one considers that when a solid is macroscopic, all that deals with its surface has little eﬀect on its macroscopic physical parameters like its pressure, its temperature. The solid will now be closed on itself (Fig. 6.3) and this will not noticeably modify its properties : one thus suppresses surface

142

Chapter 6. General Properties of the Quantum Statistics

eﬀects, of little importance for a large system. Then at a given x the properties are the same, and the wave function takes the same value, whether the particle arrives at this point directly or after having traveled one turn around before reaching x : ψ(x + Lx ) = ψ(x)

(6.47)

Lx O

x

Fig. 6.3 : The segment [0, Lx ] is closed on itself. This condition is expressed on the progressive-wave type wave function (6.39) through exp ikx (x + Lx ) = exp ikx x

(6.48)

for any x, so that the condition to be fulﬁlled is kx Lx = nx 2π 2π that is, kx = nx Lx

(6.49) (6.50)

where nx is integer. The same argument is repeated on the y and z coordinates. The wave vector k is again quantized, with basis vectors twice larger than for the condition i) above for stationary waves, the allowed values now being ã Å k = n 2π n 2π n 2π (6.51) x Lx y Ly z Lz The integers nx , ny , nz are now positive, negative or null (Fig. 6.4). Indeed the time-dependent wave function is written, to a phase, 2 k 2 1 Ψ(r, t) = √ exp i(k.r − ωt), with ω = ε = 2m Ω

(6.52)

Changing the sign of k, thus of the integers ni , produces a new wave which propagates in the opposite direction and is thus physically diﬀerent. If one of these integers is zero, this means that the wave is constant along this axis.

Free Particle in a Box ; Density of States

143

kz 2π Ly

L

x

2π

2π Lz

ky

kx Fig. 6.4: Quantization of the k-space for the periodic (B-VK) boundary conditions. Note that each dimension of the elementary box is twice that corresponding to the stationary boundary conditions (Fig. 6.2), and that now the entire space should be considered.

6.4.2

Density of States

Under these quantization conditions what is the order of magnitude of the obtained energies, and of the spacing of the allowed levels. As an example consider the framework of the periodic boundary conditions. The quantized allowed energies are given by ñ Å ã2 Å ã2 Å ã2 ô 2 2 2π 2 2π 2 2π (6.53) + ny + nz n ε= 2m x Lx Ly Lz For Lx for the order of 1 mm, a single coordinate and a mass equal to that of the proton, 2 2m

Å

2π Lx

ã2 = 13.6 eV.

ã2 Å (2π)2 0.5 × 10−10 ∼ . = 8 × 10−16 eV 1840 10−3

(6.54)

For a mass equal to that of the electron, the obtained value, 1840 times larger, is still extremely small at the electron-volt scale ! The spacing between the considered energy levels is thus very small as compared to any realistic experimental accuracy, as soon as Lx is macroscopic, that is, large with respect to atomic distances (the Bohr orbit of the hydrogen atom is associated with energies in the electron-volt range, see the above es-

144

Chapter 6. General Properties of the Quantum Statistics

timation). If the stationary boundary conditions had been chosen, the energy splitting would have been four times smaller. The practical question which is raised is : “What is the number of available quantum states in a wave vector interval d3k around k ﬁxed, or in an energy interval dε around a given ε ?” It is to answer this question that we are going to now deﬁne densities of states, which summarize the Quantum Mechanics properties of the system. It is only in a second step (§ 6.5) that Statistical Physics will come into play through the occupation factors of these accessible states. a) Wave Vector and Momentum Densities of States (B-VK Conditions) : One is estimating the number of quantum states dn of a free particle, conﬁned within a volume Ω, of wave vector between k and k + d3k, the volume d3k (2π)3 . The wave vector density of being assumed to be large as compared to Ω states Dk (k) is deﬁned by dn = Dk (k)d3k

(6.55)

We know that the allowed states are uniformly spread in the k-space and that, in the framework of the B-VK periodic boundary conditions, for each (2π)3 there is a new quantum state (Fig. 6.4) : indeed a elementary volume Ω parallelepiped has eight summits, shared between eight parallelepipeds. The searched number dn is thus equal to dn =

d3k Ω 3 = d k, (2π)3 /Ω (2π)3

(6.56)

which determines Dk (k) =

Ω (2π)3

(6.57)

The above estimation has been done from the Schroedinger equation solutions related to the space variable r (orbital solutions). If the particle carries a spin s, to a given solution correspond 2s + 1 states with diﬀerent spins, degenerate in the absence of applied magnetic ﬁeld, which multiplies the above value of the density of states [Eq. (6.57)] by (2s + 1). Thus for an electron, of spin 1/2, there are two distinct spin values for a given orbital state. For a free particle, its wave vector and momentum are related by p = k d3 p = 3 d3k

(6.58)

Free Particle in a Box ; Density of States

145

The number dn of states for a free particle of momentum vector between p and p + d3 p, conﬁned within the volume Ω, is written, from eq. (6.56), dn =

p Ω Ω d3 = 3 d3 p = Dp ( p)d3 p 3 3 (2π) h

(6.59)

Here the (three-dimensional) momentum density of states has been deﬁned through Dp ( p) =

Ω h3

(6.60)

Note that the distance occupied by a quantum state, on the momentum axis of the phase space corresponding to the one-dimension motion of a single h . This is consistent with the fact that a cell of this phase space particle, is Lx has the area h and that here ∆x = Lx , since, the wave function being of plane-wave type, it is delocalized on the whole accessible distance. b) Three-Dimensional Energy Density of States (B-VK Conditions) : The energy of a free particle only contains the kinetic term inside the allowed volume, it only depends of its momentum or wave vector modulus : ε=

2 k 2 p2 = 2m 2m

(6.61)

All the states corresponding to a given … energy ε are, in the k-space, at the 2mε surface of a sphere of radius k = . If one does not account for the 2 particle spin, the number of allowed states between the energies ε and ε + dε is the number of allowed values of k between the spheres of radii k and k + dk, 2 kdk. (For clarity in the sketch, Fig. dk being related to dε through dε = m 6.5 is drawn in the case of a two-dimensional problem.) The concerned volume is 4πk 2 dk, the number dn of the searched states, without considering the spin variable, is equal to Ω Ω 4πk 2 dk = 4πk.kdk (2π)3 (2π)3 … Ω 2mε m = 2 dε 2π 2 2

(6.62) (6.63)

Thus a density of states in energy is obtained, deﬁned by dn(ε) = D(ε) dε, with Å ã √ (2s + 1) 2m 3/2 (6.64) D(ε) = CΩ ε with C = 4π 2 2

146

Chapter 6. General Properties of the Quantum Statistics

ky ε+ ε

dε

kx O

2π Lx

2π Ly Fig. 6.5: Number of states of energy between ε and ε + dε, (two-dimensional problem). This expression of the energy density of states, which includes the spin degeneracy, is valid for any positive energy ; the density of states vanishes for ε < 0, since then no state can exist, the energy being only of kinetic origin for the considered free particle. c) Densities of States (Stationary-Waves Conditions) We have seen that, in the case of stationary-wave conditions, all the physically distinct solutions are obtained when restricting the wave vectors choice to those with all three positive components. The volume corresponding to the states of energies between ε and ε + dε must thus be limited to the corresponding trihedral of the k-space. This is the 1/8th of the space between the two 1 spheres of radii k and k + dk, i.e., 4πk 2 dk, dk being still related to dε by 8 dε =

2 kdk m

(6.65)

However the wave vector density of states is eight times larger for the stationary-wave quantization condition than for the B-VK condition (the allowed points of the k-space are twice closer in the x, y and z directions). Then one obtains a number of orbital states, without including the spin degeneracy, of dn =

Ω1 4πk 2 dk π3 8

(6.66)

Fermi-Dirac Distribution ; Bose-Einstein Distribution

147

The density D(ε) thus has exactly the same value than in b) when the spin degeneracy (2s + 1) is introduced. One sees that both types of quantization condition are equivalent for a macroscopic system. In most of the remaining parts of this course, the B-VK conditions will be used, as they are simpler because they consider the whole wave vector space. Obviously, in speciﬁc physical conditions, one may have to use the stationary-wave quantization conditions. Notes : In presence of a magnetic potential energy, which does not aﬀect the orbital part of the wave function, the density of states Dk (k) is unchanged. On the contrary, D(ε) is shifted (see § 7.1.3). As an exercise, we now calculate the energy density of states for free particles in the two- or one-dimension space. The relations between ε and k, between dε and dk are not modiﬁed [(6.61) and (6.65)]. – The two-dimension elementary area is 2πkdk, the k -density of states Lx Ly is given by without taking the spin in account and the energy (2π)2 density of states is a constant. – In one dimension, the elementary length is 2dkx (kx and −kx correspond L to the same energy, whence the factor 2), the kx -density of states is 2π 1 without spin, and D(ε) varies like √ . D(ε) diverges for ε tending to ε zero but its integral, which gives the total number of accessible states from ε = 0 to ε = ε0 , is ﬁnite.

6.5

Fermi-Dirac Distribution ; Bose-Einstein Distribution

We have just seen that, for a macroscopic system, the accessible energy levels are extremely close and have deﬁned a density of states, which is a discrete function of energy ε : in the large volume limit, it is as if D(ε) was a function of the continuous variable ε. Let us now return to Statistical Physics in the framework of the large volumes limit. From the average number of occupation nk (see § 6.3) of an energy level εk , one deﬁnes the energy function “ occupation factor” or “distribution,” of Fermi-Dirac or of Bose-Einstein according to the nature of the considered

148

Chapter 6. General Properties of the Quantum Statistics

particles : 1

fF D (ε) =

eβ(ε−µ)

+1

=

1 eβε−α

+1

1 1 = eβ(ε−µ) − 1 eβε−α − 1

fBE (ε) =

(6.67) (6.68)

Recall that α and the chemical potential µ are related by α = βµ with 1 β= . kB T The consequences of the expression of their distribution will be studied in details in chapter 7 for free fermions and in chapter 9 for bosons.

6.6

Average Values of Physical Parameters at T in the Large Volumes Limit

On the one hand, the energy density of states D(ε) summarizes the Quantum Mechanical properties of the system, i.e., it expresses the solutions of the Schroedinger equation for an individual particle. On the other hand, the distribution f (ε) expresses the statistical properties of the considered particles. To calculate the average value of a given physical parameter, one writes that in the energy interval between ε and ε + dε, the number of accessible states is D(ε)dε, and that at temperature T and for a parameter α or a chemical potential µ, these levels are occupied according to the distribution f (ε) (of Fermi-Dirac or Bose-Einstein). Thus +∞ D(ε)f (ε) dε (6.69) N= −∞ +∞

U=

εD(ε)f (ε) dε

(6.70)

−∞

The temperature and the chemical potential are included inside the distribution expression. In fact the ﬁrst equation allows one to determine the chemical potential value. To obtain the grand potential, one generalizes to a large system relations (6.31) and (6.32) that were established for discrete parameters : +∞ A = +kB T D(ε) ln(1 − f (ε)) dε for fermions (6.71) −∞ +∞

A = −kB T

D(ε) ln(1 + f (ε)) dε for bosons −∞

(6.72)

Common Limit of the Quantum Statistics

6.7 6.7.1

149

Common Limit of the Quantum Statistics Chemical Potential of the Ideal Gas

In the case of free particles (fermions or bosons) of spin s, conﬁned in a volume Ω, one expresses the number of particles (6.69) using the expression (6.64) of the density of states D(ε) : √ εdε (2s + 1)Ω(2m)3/2 +∞ (6.73) N= 2 3 βε−α 4π e ∓1 0 where, in the denominator, the sign − is for bosons and the sign + for fermions. If the temperature is ﬁnite, the dimensionless variable of the integral is βε = x, so that this relation can be written : √ ã Å N 2mkB T 3/2 (2s + 1) +∞ xdx = (6.74) x−α ∓ 1 Ω 2 4π 2 e 0 The physical parameters are then put into the left member, which depends of the system density and temperature as √ Å ã3/2 h2 1 1 N 4π 2 N π 1 N (λth )3 = constant λ3th = Ω (2π)3 2s + 1 23/2 mkB T Ω 2 2s + 1 Ω (6.75) The thermal de Broglie wavelength λth appears, as deﬁned in § 4.3. The quantity (6.75) must be equal to the right member integral, which is a function of α such that ∞ √ xdx (6.76) u∓ (α) = x−α e ∓1 0 √ In fact (6.75) expresses, to the factor π/2(2s+1), the cube of the ratio of the thermal de Broglie wavelength λth to the average distance between particles d = (Ω/N )1/3 . It is this comparison which, in § 4.3, had allowed us to deﬁne the limit of the classical ideal gas model. Since (6.75) and (6.76) are equal, one obtains : √ 1 πN (λth )3 (6.77) u∓ (α) = 2s + 1 2 Ω The integral u∓ (α) can be numerically calculated versus α, it is a monotonous increasing function of α for either Fermi-Dirac (F D) or Bose-Einstein (BE) Quantum Statistics (Fig. 6.6). We already know the sign or the value of u(α) for some speciﬁc physical situations : It was noted in § 6.2.2 that the expression of f (ε) for Bose-Einstein statistics implies that α < βε1 , i.e., α < 0 for free bosons, since then ε1 is zero. We

150

Chapter 6. General Properties of the Quantum Statistics

uF D

u± (α)

u+ (α) = uF D u− (α) = uBE uBE constant classical limit αBE

αF D O

N 3 λ Ω th α

Fig. 6.6: Sketch of the function u± (α) and determination of α from the physical data of the considered system.

will see in chapter 9 of this book that, in some situations, α vanishes, thus producing the Bose-Einstein condensation, which is a phase transition. In the case of free fermions at low temperature, it will be seen in the next chapter that µ is larger than ε1 , which is equal to zero, i.e., α > 0. When α tends to minus inﬁnity, the denominator term takes the same expression for both fermions and bosons. The integration is exactly performed : u∓ (α) e

α

+∞

0

√ √ −x α π

1 xe dx = e 2

(6.78)

Considering the expression of the right member of 6.76, this limit does occur if the particle density n = N/Ω tends to zero or the temperature tends to inﬁnity. In the framework of this classical limit, one deduces eα =

N λ3th Ω (2s + 1)

(6.79)

which allows one to express the chemical potential µ, since α = µ/kB T . One

Common Limit of the Quantum Statistics

151

ﬁnds, if the spin degeneracy is neglected, i.e., if one takes 2s + 1 = 1, ñ Å ã ô Ω 2πmkB T 3/2 µ = −kB T ln (6.80) N h2 the very expression obtained in § 4.4.3 for the ideal gas.

6.7.2

Grand Canonical Partition Function of the Ideal Gas

Let us look for the expression of A in the limit where α tends to minus inﬁnity. For both fermions and bosons, one gets the same limit, by developing the Naperian logarithm to ﬁrst order : it is equal to eα−βεk (6.81) A = −kB T k

e−βεk

(6.82)

One recognizes the one-particle partition function z1 = e−βεk

(6.83)

= −kB T e

α

k

k

(its value for the ideal gas was calculated in § 4.4.1), whence A = −kB T eα z1

(6.84)

A = −kB T ln ZG ,

(6.85)

By deﬁnition

thus in the common limit of both Quantum Statistics, i.e., in the classical limit conditions, ZG is equal to ZG = exp(eα z1 ) =

N

eαN

z1N N!

(6.86)

Now the general deﬁnition of the grand canonical partition function ZG , using the N -particle canonical partition function ZcN , is ZG = eαN ZcN (6.87) N

This shows that in the case of identical particles, for sake of consistency, the N -particles canonical partition function must contain the factor 1/N !, the

152

Chapter 6. General Properties of the Quantum Statistics

“memory” of the Quantum Statistics, to take into account the indistinguishability of the particles : ZcN =

1 N z N! 1

(6.88)

This factor ensures the extensivity of the thermodynamical functions, like S, and allows to solve the Gibbs paradox (there is no extra entropy when two volumes of gases of the same density and the same nature are allowed to mix into the same container.) (see § 4.4.3). Notes : The above reasoning allows one to grasp how, in the special case of free particles, one goes from a quantum regime to a classical one. We did not specify to which situation corresponds the room temperature, the most frequent experimental condition. We will see that for metals, for example, it corresponds to a very low temperature regime, of quantum nature (see chapter 7).

Summary of Chapter 6 To account for the occupation conditions of the energy states as stated by the Pauli principle, it is simpler to work in the grand canonical ensemble, in which the number of particles is only given in average : then, at the end of the calculation, the value of the chemical potential is adjusted so that this average number coincides with the real number of particles in the system. As soon as the system is macroscopic, the ﬂuctuations around the average particle number predicted in the grand canonical ensemble are much smaller than the resolution of any measure. The grand partition function gets factorized by energy level and takes a different expression for fermions and for bosons. Remember the distribution, which provides the average number of particles, of chemical potential equal to µ, on the energy level ε at temperature T = 1/kB β : for fermions fF D (ε) =

1 exp β(ε − µ) + 1

(6.89)

fBE (ε) =

1 exp β(ε − µ) − 1

(6.90)

for bosons

For a free particle in a box, we calculated the location of the allowed energy levels. The periodic boundary conditions of Born-Von Kármán (or in an analogous way the stationary-waves conditions) evidence quantized levels, and for a box of macroscopic dimensions (“large volume limit”), the allowed levels are extremely close. One then deﬁnes an energy density of states such that the number of allowed states between ε and ε + dε is equal to dn = D(ε)dε For a free particle in the three-dimensional space D(ε) ∝ 153

(6.91) √

ε.

154

Summary of Chapter 6

In the large volume limit the average values at temperature T of physical parameters are expressed as integrals, which account for the existence of allowed levels through the energy density of states and for their occupation by the particles at this temperature through the distribution term : +∞ D(ε)f (ε) dε (6.92) N= −∞ +∞

U=

εD(ε)f (ε) dε

(6.93)

−∞

We showed how for free particles the Classical Statistics of the ideal gas is the common limit of the Fermi-Dirac and Bose-Einstein statistics when the temperature is very high and/or the density very low. The factor 1/N ! in the canonical partition function of the ideal gas is then the “memory” of the Quantum Statistics properties.

Chapter 7

Free Fermions Properties We now come to the detailed study of the properties of the Fermi-Dirac quantum statistics. First at zero temperature we will study the behavior of free fermions, which shows the consequences of the Pauli principle. Then we will be concerned by electrons, considered as free, and will analyze two of their properties which are temperature-dependent, i.e., their speciﬁc heat and the “ thermionic” emission. In the next chapter we will understand why this free electron model indeed properly describes the properties of the conduction electrons in metals. In fact, in the present chapter we follow the historical order : the properties of electrons in metals, described by this model, were published by Arnold Sommerfeld (electronic speciﬁc heat, Spring 1927) and Wolfang Pauli (electrons paramagnetism, December 1926), shortly after the statement of the Pauli principle at the beginning of 1925.

7.1

7.1.1

Properties of Fermions at Zero Temperature Fermi Distribution, Fermi Energy

The Fermi-Dirac distribution, presented in § 6.5, gives the average value of the occupation number of the one-particle state of energy ε, and is a function of the chemical potential µ. Its expression is ï ò 1 β(ε − µ) 1 f (ε) = β(ε−µ) = 1 − tanh (7.1) 2 2 e +1 155

156

Chapter 7. Free Fermions Properties

1 It is plotted in Fig. 7.1 and takes the value for ε = µ for any value of β and 2 thus of the temperature. This particular point is a center of inversion of the curve, so that the probability of “nonoccupation” for µ− δε, i.e., 1 − f (µ− δε), has the same value as the occupation probability for µ + δε, i.e., f (µ + δε). β At the point ε = µ the tangent slope is − : the lower the temperature, the 4 larger this slope. f 1−f 1

2 kB T

1 − f (ε)

f (ε) 0

µ

ε

Fig. 7.1 : Variation versus energy of f (ε) and 1 − f (ε). In particular when T tends to zero, f (ε) tends to a “step function” (Heaviside function). One usually calls Fermi energy the limit of µ when T tends to zero ; it is noted εF , its existence and its value will be speciﬁed below. Then one obtains at T = 0 : ® f (ε) = 1 if ε < εF (7.2) f (ε) = 0 if ε > εF At T = 0 K the ﬁlled states thus begin at the lowest energies, and end at εF , the last occupied state. The states above εF are empty. This is in agreement with the Pauli principle : indeed in the fundamental state the particles are in their state of minimum energy but two fermions from the same physical system cannot occupy the same quantum state, whence the requirement to climb up in energy until the particles all located. In the framework of the large volume limit, to which this course is limited, the total number N of particles in the system satisﬁes ∞ D(ε)f (ε)dε (7.3) N= 0

where D(ε) in the one-particle density of states in energy. This condition determines the µ value.

Properties of Fermions at Zero Temperature

157

At zero temperature, the chemical potential, then called the “Fermi energy εF ,”, is thus deduced from the condition that εF N= D(ε)dε (7.4) 0

The number N corresponds to the hatched area of Fig. 7.2.

D(ε)

f (ε)

T = OK

0

εF

ε

Fig. 7.2: Density of states D(ε) of three-dimensional free particles and determination of the Fermi energy at zero temperature. The hatched area under the curve provides the number N of particles in the system. For a three-dimensional system of free particles, with a spin s = 1/2, from (6.64) √ Ω (2m)3/2 ε 2π 2 3 D(ε) = 0

D(ε) =

for ε > 0 for

(7.5)

ε≤0

The integral (7.4) is then equal to N=

2 Ω 3/2 (2m)3/2 εF 3 2π 2 3

(7.6)

Å ã N 2/3 2 3π 2 2m Ω

(7.7)

that is, εF =

This expression leads to several comments :

158

Chapter 7. Free Fermions Properties

– N and Ω are extensive, thus εF is intensive ; – εF only depends on the density N/Ω ; – the factor of 2 /2m in (7.7) is equal to kF2 , where kF is the Fermi wave vector : in the k-space the surface corresponding to the constant energy εF , called “Fermi surface,” contains all the ﬁlled states. In the case of free particles, it is a sphere, the “Fermi sphere,” of radius kF . One can √ also deﬁne the momentum pF = kF = 2mεF . Then expression (7.6) can be obviously expressed versus kF or pF for spin-1/2 particles : N=

2Ω 4 3 2Ω 4π 3 πk = 3 p (2π)3 3 F h 3 F

(7.8)

p)), the one-particle density The ﬁrst factor is equal to Dk (k) (resp. Dp ( of states in the wave vector- (resp. momentum-) space. Expression (7.8) is also directly obtained by writing : N= 0

kF

Dk (k)d3k .

– εF corresponds, for metals, to energies of the order of a few eV. Indeed, let us assume now (see next chapter) that the only electrons to be considered are those of the last incomplete shell, which participate in the conduction. Take the example of copper, of atomic mass 63.5 g and density 9 g/cm3 . There are thus 9N /63.5 = 0.14N copper atoms per cm3 , N being the Avogadro number. Since one electron per copper atom participates in conductivity, the number of such electrons per cm3 is of 0.14N = 8.4 × 1022 /cm3 i.e., 8.4 × 1028 / m3 . This corresponds to a wave vector kF = 1.4 × 1010 m−1 , and to a velocity at the Fermi level kF of 1.5 × 106 m/sec. One deduces εF = 7 eV. vF = m Remember that the thermal energy at 300 K, kB T , is close to 25 meV, so that the Fermi energy εF corresponds to a temperature TF such that kB TF = εF , in the 104 or 105 K range according to the metal (for copper 80, 000 K) !

7.1.2

Internal Energy and Pressure of a Fermi Gas at Zero Temperature

The internal energy at zero temperature is given by εF εD(ε)dε U= 0

(7.9)

Properties of Fermions at Zero Temperature

159

For the three-dimensional density of states (7.5), U is equal to : U=

2 Ω 5/2 (2m)3/2 εF 5 2π 2 3

(7.10)

i.e., by expressing N using (7.6), U= Besides,

3 N εF 5 Å

∂U P =− ∂Ω

(7.11)

ã (7.12) N

From (7.10) and (7.7), U varies like N 5/3 Ω−2/3 , so that at constant particles number N P =

2U 3Ω

(7.13)

For copper, one obtains P = 3.8 × 1010 N/m2 = 3.7 × 105 atmospheres. This very large pressure, exerted by the particles on the walls of the solid which contains them, exists even at zero temperature. Although the state equation (7.13) is identical in the cases of fermions and of the ideal gas, the two situations are very diﬀerent : for a classical gas, when the temperature tends to zero, both the internal energy and the pressure vanish. For fermions, the Pauli principle dictates that levels of nonzero energies are ﬁlled. The fact that electrons in the same orbital state cannot be more than a pair, and of opposite spins, leads to their mutual repulsion and to a large average momentum, at the origin of this high pressure.

7.1.3

Magnetic Properties. Pauli Paramagnetism

A gas of electrons, each of them carrying a spin magnetic moment equal to If N+ the Bohr magneton µ B , is submitted to an external magnetic ﬁeld B. equal to +µB , is the number of electrons with a moment projection along B N− that of moment −µB , the gas total magnetic moment (or magnetization) is along B M = µB (N+ − N− )

(7.14)

The density of states Dk (k) is the same as for free electrons. Since the potential energy of an electron depends on the direction of its magnetic moment, the total energy of an electron is ε± =

2k 2 ∓ µB B 2m

(7.15)

160

Chapter 7. Free Fermions Properties

The densities of states D+ (ε) and D− (ε) now diﬀer for the two magnetic moment orientations, whereas in the absence of magnetic ﬁeld, D+ (ε) = D− (ε) = D(ε)/2 : here ® D+ (ε) = 12 D(ε + µB B) (7.16) D− (ε) = 12 D(ε − µB B) Besides, the populations of electrons with the two orientations of magnetic moment are in equilibrium, thus they have the same chemical potential. The solution of this problem is a standard exercise of Statistical Physics. Here we just give the result, in order to comment upon it : one obtains the magnetization M = µB · µB B · D(εF )

(7.17)

This expresses that the moments with the opposite orientations compensate, except for a small energy range of width µB B around the Fermi energy. The 1 M corresponding (Pauli) susceptibility, deﬁned by lim , is equal to µ0 Ω B B→0 χ=

1 3 N µ2B kB T 1 3 N µ2B · = µ0 2 Ω εF µ0 2 Ω kB T εF

(7.18)

with µ0 = 4π × 10−7 . It is reduced with respect to the classical result (Curie law) by a factor of the order of kB T /εF , which expresses that only a very small fraction of the electrons participate in the eﬀect (in other words, only a few electrons can reverse their magnetic moment to have it directed along the external ﬁeld).

7.2

7.2.1

Properties of Fermions at Non-Zero Temperature Temperature Ranges and Chemical Potential Variation

At nonvanishing temperature, one again obtains the chemical potential µ position from expression (7.3), which links µ to the total number of particles : ∞ D(ε)f (ε)dε N= 0 ∞ 1 = D(ε) β(ε−µ) dε (7.3) e +1 0

Properties of Fermions at Non-Zero Temperature

161

D(ε) f (ε) T TF

f (ε)D(ε)

0

µ

εF

ε

Fig. 7.3 : Determination of the chemical potential µ for T TF .

Relation (7.3) implies that µ varies with T . Fig. 7.3 qualitatively shows for kB T εF the properties which will be calculated in Appendix 7.1. Indeed the area under the curve representing the product D(ε)f (ε) gives the number of particles. At T = 0 K, one has to consider the area under D(ε) up to the energy εF ; at small T , a few states above µ are ﬁlled and an equivalent number below µ are empty. Since the function D(ε) is increasing, to maintain an area equal to N, µ has to be smaller than εF , of an amount proportional to D (µ). The exact relation allowing one to deduce µ(T ) for the two Quantum Statistics was given in (6.74) : recalling that α = βµ, and substituting (7.5) into (7.3) one obtains for fermions √ ∞ Å ã3/2 xdx 2 2 N = 2π · (7.19) exp(x − α) + 1 Ω 2mkB T 0 Å ã 2 εF 3/2 = (7.20) 3 kB T In the integral we deﬁned βε = x. Thus α is related to the quantity εF /kB T = TF /T . – When T vanishes, the integral (7.19) tends to inﬁnity : the exponential in the denominator no longer plays any role since α tends to inﬁnity (εF is ﬁnite and β tends to inﬁnity). The right member also tends to

162

Chapter 7. Free Fermions Properties

inﬁnity. One thus has to return to equation (7.4). – We have just studied the low temperature regime T TF , which takes place in metals at room temperature (TF ∼ 104 K) and corresponds to Fig. 7.3. In astrophysics, the “white dwarfs” enter in this framework. They are stars at the end of their existence, which have already burnt all their nuclear fuel and therefore have shrunk. Their size is of the order of a planet (radius of about 5000 km) for a mass of 2 × 1030 kg (of the order of that of the sun), their temperature T of 107 K. Their high density in electrons determines εF . TF being in the 109 K range, the low temperature regime is achieved ! This regime for which the Quantum Statistics apply is called the “degenerate” (in opposition to “classical”) regime : as already mentioned about Pauli paramagnetism, and will be seen again below, the “active” particles all have an energy close to µ, this is almost as if this level was degenerate in the Quantum Mechanics meaning. – On the other hand, when T TF , the integral (7.19) vanishes. This is realized either at low density (εF small), or at given density when T tends to inﬁnity. Then e−α tends to inﬁnity, and since α = βµ with β tending to zero, this implies that µ tends to −∞. One then ﬁnds the Maxwell-Boltzmann Classical Statistics, as discussed in § 6.7.2. – Between these two extreme situations, µ thus decreases from εF to −∞ when the ratio T /TF increases. When studying systems at room temperature, the regime is degenerate in most cases. For properties mainly depending on the Pauli principle, one can consider then to be at zero temperature. This is the situation when ﬁlling the electronic shells in atoms according to the “Aufbau (construction) principle”, widely used in Chemistry courses : the Quantum Mechanics problem is ﬁrst solved in order to ﬁnd the one-electron energy levels of the considered atom. The solutions ranked in increasing energies are the states 1s, 2s, 2p, 3s, 3p, 4s, 3d, etc. (except some particular cases for some transition metals). These levels are then ﬁlled with electrons, shell after shell, taking into account the level degeneracy (K shell, n = 1, containing at most two electrons, L shell of n = 2, with at most eight electrons, etc.). Since the energy splittings between levels are of the order of a keV for the deep shells and of an eV for the valence shell, the thermal energy remains negligible with respect to the energies coming into play and it is as if the temperature were zero. On the other hand, if one is interested in temperature-dependent properties, like the speciﬁc heat or the entropy, which are derivatives versus temperature of thermodynamical functions, more exact expressions are required and at low temperature one will proceed through limited developments versus the variable T /TF .

Properties of Fermions at Non-Zero Temperature

7.2.2

163

Speciﬁc Heat of Fermions

The very small contribution of electrons to a metal speciﬁc heat, which follow the Dulong and Petit law (see Ch. 2, § (2.4.5)) only introducing the vibrations of the lattice atoms, presented a theoretical challenge in the 1910s. This lead to a questioning of the electron conduction model, as proposed by Paul Drude in 1900. Å ã ∂U of a fermions assembly, one To calculate the speciﬁc heat Cv = ∂T N,Ω needs to express the temperature dependence of the internal energy. One can already make an estimate of Cv from the observation of the shape of the Fermi distribution : the only electrons which can absorb energy are those occupying the states in a region of a few kB T width around the chemical potential, and each of them will increase its energy by an amount of the order kB T to reach an empty state. The number of concerned electrons is thus ∼ kB T · D(µ) and for a threekB T dimensional system at low temperature (

1, so that µ and εF are εF close), µ N∼ D(ε)dε ∝ µ3/2 , (7.21) 0

3N dN = = D(µ) i.e., dµ 2µ

(7.22)

The total absorbed energy is thus of the order of Å ã 3N U ∼ kB T · · kB T 2 µ Consequently, Cv =

dU ∼ N kB dT

Å

kB T µ

(7.23)

ã (7.24)

These considerations evidence a reduction of the value of Cv with respect to the classical value N kB in a ratio of the order kB T /µ 1, because of the very small number of concerned electrons. The exact approach refers to a general method of calculation of the integrals in which the Fermi distribution enters : the Sommerfeld development. The complete calculation is given in Appendix 7.1. Here we just indicate its principle. At a temperature T small with respect to TF , one looks for a limited development in T /TF of the fermions internal energy U , which depends of T , and of

164

Chapter 7. Free Fermions Properties

the chemical Å ã potential, itself temperature-dependent. The searched quantity dU is then equal to Cv = dT N,Ω Cv =

3 π2 N kB · 2 3

Å

kB T εF

ã (7.25)

which is indeed of the form (7.24), as far as µ is not very diﬀerent from εF (see Appendix 7.1). The ﬁrst factor in (7.25) is the speciﬁc heat of a monoatomic ideal gas ; the last factor is much smaller than unity in the development conditions, it ranges between 10−3 and 10−2 at 300 K (for copper it is 3.6 × 10−3 ) and expresses the small number of eﬀective electrons. The electron’s speciﬁc heat in metals is measured in experiments performed at temperatures in the kelvin range. In fact the measured total speciﬁc heat includes the contributions of both the electrons and the lattice Cvtotal = Cvel + Cvlattice

(7.26)

The contribution Cvlattice of the lattice vibrations to the solid speciﬁc heat was studied in Ch. 2, § 2.4.5. The Debye model was mentioned, which correctly interprets the low temperature results by predicting that Cvlattice is proportional to T 3 , as observed in non-metallic materials. The lower the temperature, the less this contribution will overcome that of the electrons. The predicted variation versus T for Cvtotal is thus given by Cvtotal = γT + aT 3

(7.27)

Cvtotal

= γ + aT 2

(7.28)

π 2 kB 2 εF

(7.29)

T with, from (7.25),

γ = N kB

From the measurement of γ one deduces a value of εF of 5.3 eV for copper whereas in § 7.1.1 we calculated 7 eV from the value of the electrons concentration, using (7.7). In the case of silver the agreement is much better since the experiment gives εF = 5.8 eV, close to the prediction of (7.7). The possible discrepancies arise from the fact that the mass m used in the theoretical expression of the Fermi energy was that of the free electron. Now a conduction electron is submitted to interactions with the lattice ions, with the solid vibrations and the other electrons. Consequently, its motion is altered,

Properties of Fermions at Non-Zero Temperature

165

which can be expressed by attributing it an eﬀective mass m∗ . This concept will be explained in § 8.2.3 and § 8.3.3. Deﬁning εF =

ã2/3 Å 2 2N , 3π 2m∗ V

(7.30)

one deduces from these experiments a mass m∗ = 1.32m for copper, m∗ = 1.1m for silver, where m is the free electron mass. Note that the fermion entropy can be easily deduced from the development of the Appendix : we saw that A = −P Ω (see §Å3.5.3), ã so that, according to 2 ∂A (7.13), A = − U . Besides, from (3.5.3), S = − . One deduces 3 ∂T µ,Ω S = N kB

π 2 kB T = Cv 2 εF

(7.31)

The electrons entropy thus satisﬁes the third law of Thermodynamics since it vanishes with T .

7.2.3

Thermionic Emission ε

0

V εF

f (ε) 1

0

Fig. 7.4: Potential well for the electrons of a metal. In order to extract an electron at zero temperature, one illuminates the solid with a photon of energy higher than V − εF . The walls of the “potential box” which conﬁnes the electrons in a solid are not really inﬁnitely high. The photoelectrical eﬀect, evidenced by Heinrich

166

Chapter 7. Free Fermions Properties

Hertz (1887), and interpreted by Albert Einstein (this was the argument for Einstein’s Nobel prize in 1921), allows one to extract electrons from a solid, at room temperature, by illumination with a light of high enough energy : for zinc, a radiation in the near ultraviolet is required (Fig. 7.4). The energy’s origin is here the state of a free electron with zero kinetic energy, i.e., with such a convention the Fermi level has a negative energy. The potential barrier is equal to V . To extract an electron at zero temperature, a photon must bring it at least V − εF , an energy usually in the ultraviolet range. When heating this metal to a high temperature T , the Fermi distribution is spreading toward little negative or even positive energies, so that a few electrons are then able to pass the potential barrier and become free (Fig. 7.5). One utilizes tungsten heated to about 3000 K to make the ﬁlaments of the electrons cannons of TV tubes or electronic microscopes, as its fusion temperature is particularly high (3700 K). ε

0

V µ

f (ε)

1/2

Fig. 7.5: Filling of the electronic states in a metal heated to a high temperature T . One way to calculate the thermal current consists in expressing that a particle escapes the metal only if it can overcome the potential barrier, i.e. if its kinetic energy in the direction perpendicular to the surface, taken as z axis, is large enough. It should satisfy p2 p2 ≥ z ≥V (7.32) 2m 2m We know that the electronic energies, and thus V and V − µ, are of the order of a few eV : whatever the ﬁlament temperature, the system is in a regime

Properties of Fermions at Non-Zero Temperature

167

where kB T remains small with respect to the concerned energies, and where kB T

1. Consequently, µ is still close to εF , and in the Fermi distribution εF f (ε) = exp β

Ä

1 p2 2m

ä −µ +1

(7.33)

the denominator exponent is large for the electrons likely to be emitted. The Fermi distribution can then be approximated by a Maxwell-Boltzmann distribution p2

f (ε) eβµ e−β 2m

(7.34)

There remains to estimate the number ∆n of electrons moving toward the surface ∆S during a time ∆t. This is a standard problem of kinetic theory of gases (the detailed method of calculation is analogous to the kinetic calculation of the pressure, § 4.2.2). One obtains a total current jz = ∆n/∆S∆t equal to jz =

2 ∆n = 3 2πm(kB T )2 e−β(V −µ) ∆S∆t h

(7.35)

This expression shows the very fast variation of the thermal emission current with the temperature T , in T 2 e−(V −µ)/kB T . It predicts that a 0.3 × 0.3 mm2 tungsten tip emits a current of 15 mA at 3000 K and of only 0.8µA at 2000 K. Practically, to obtain a permanent emission current, one must apply an external electric ﬁeld : in the absence of such a ﬁeld, the ﬁrst electrons emitted into vacuum create a space charge, associated with an antagonistic electric ﬁeld, which prevents any further emission. Another way of calculating the emitted current consists in considering that in the TV tube, in the absence of applied electrical potential, an equilibrium sets up at temperature T between the tungsten electrons which try to escape the metal and those outside the metal that try to enter it. The equilibrium between electrons in both phases is expressed through the equality of their chemical potentials. Now, outside the solid the electron concentration is very small, so that one deals with a classical gas which follows the Maxwell-Boltzmann statistics. The electron current toward the solid is calculated like for an ideal gas. When an electric ﬁeld is applied, this current vanishes but the emission current, which is its opposite, remains. Note that, in order to be rigorous, all these calculations should take into account the energy levels of the metal and its surface, which would be very

168

Chapter 7. Free Fermions Properties

complicated ! Anyhow, the above estimates qualitatively describe the eﬀect and lead to correct orders of magnitude. On this example of the thermionic emission we have shown a practical application, in which electrons follow the high temperature limit of the Fermi-Dirac statistics.

Summary of Chapter 7 The properties of free fermions are governed by the Fermi-Dirac distribution : fF D (ε) =

1 exp β(ε − µ) + 1

At zero temperature the fermions occupy all the levels of energy lower or equal to εF , the Fermi energy, which is deﬁned as the chemical potential µ for T = 0 K. The number N of particles of the system and the Fermi energy are linked at T = 0 K through εF N= D(ε)dε 0

This situation, related to the Pauli principle, induces large values of the internal energy and of the pressure of a system of N fermions. An excitation, of thermal or magnetic origin, of a fermions system, only aﬀects the states of energy close to µ or εF . Consequently, the fermions’ magnetic susceptibility and speciﬁc heat are reduced with respect to their values in a classical system. In the case of the speciﬁc heat, the reduction is a ratio of the order of kB T /µ, where µ is the chemical potential, of energy close to εF . The magnetic susceptibility and the speciﬁc heat of electrons in metals are well described in the framework of the free electrons model developed in this chapter.

169

Appendix 7.1 Calculation of Cv Through the Sommerfeld Development ã dU of fermions is deduced from the expression dT N,Ω of the internal energy U (T ) versus temperature. Å

The speciﬁc heat Cv =

We have therefore to calculate, at nonvanishing temperature, ∞ f (ε)D(ε)dε N= 0 ∞ U= εf (ε)D(ε)dε

(7.36) (7.37)

0

that is, more generally, g(ε) =

+∞

+∞

g(ε)D(ε)f (ε)dε =

ϕ(ε)f (ε)dε

−∞

(7.38)

−∞

where ϕ(ε) is assumed to vary like a power law of ε for ε > εminimum, to be null for smaller values of ε. We know that, owing to the shape of the Fermi distribution, εF lim g(ε) = ϕ(ε)dε T →0

We write the diﬀerence

+∞

∆= −∞

(7.39)

−∞

ϕ(ε)f (ε)dε −

µ

ϕ(ε)dε

(7.40)

−∞

which is small when kB T µ (βµ 1). This diﬀerence introduces ϕ (µ) : if ϕ (µ) is zero, because of the symmetry of f (ε) around µ, the contributions on 171

172

Appendix 7.1

both sides of µ compensate. Besides, we understood in § 7.2.1 that, to maintain the same N when the temperature increases, µ gets closer to the origin, for a three-dimensional system for which ϕ (µ) > 0 when D(ε) is diﬀerent from √ zero (ϕ(ε) ∝ ε). Let us write ∆ explicitly : µ ∆= dεϕ(ε)f (ε) − −∞ µ

= −∞

1 eβ(ε−µ)

∞

dεϕ(ε) +

−∞

Å dεϕ(ε)

µ

+1

dεϕ(ε)f (ε)

(7.41)

µ

ã −1 +

∞

dεϕ(ε)

µ

1 eβ(ε−µ)

+1

(7.42)

Here a symmetrical part is played by either the empty states below µ [the probability for a state to be empty is 1 − f (ε)], or the ﬁlled states above µ, all in small numbers if the temperature is low. We take β(ε − µ) = x, βdε = dx ∞ 0 Å Å ã ã x x dx dx 1 1 ∆=− ϕ µ+ ϕ µ+ + −x β β 1 + e β β 1 + ex −∞ 0 ∞ ∞ Å Å ã ã x 1 x 1 dx dx ϕ µ− ϕ µ+ + =− x β β 1 + e β β 1 + ex ∞0 ï Å ã Å 0 ãò x x dx 1 ϕ µ+ − ϕ µ− = x β 1 + e β β 0 ñ ô Å ã 1 ∞ dx x ϕ(3) (µ) x 3 = · 2 ϕ (µ) + + ... β 0 1 + ex β 3! β =

∞ ϕ(2n+1) (µ) In β 2n+2 n=0

(7.43) (7.44) (7.45) (7.46) (7.47)

with In =

2 (2n + 1)!

∞

dx 0

x2n+1 ex + 1

(7.48)

The integrals In are deduced from the Riemann ζ functions (see the section “A few useful formulae”). In particular at low temperature the principal term π2 of ∆ is equal to (kB T )2 ϕ (µ). 6 Let us return to the calculation of the speciﬁc heat Cv . For the internal energy of nonrelativistic particles, ϕ(ε) = ε · Kε1/2 = Kε3/2 for ε > 0, ϕ(ε) = 0 for ε < 0. Then µ π2 U (T ) = (kB T )2 (Kε3/2 )µ + O(T 4 ) Kε3/2 dε + (7.49) 6 0

Appendix 7.1

173

but µ depends on T since µ εF π2 1/2 2 1/2 4 N= (kB T ) (Kε )µ + O(T ) = Kε dε + Kε1/2 dε 6 0 0

(7.50)

expresses the conservation of the particles number. One deduces from (7.50) : ñ µ(T ) = εF

π2 1− 12

Å

kB T εF

ã2 ô + O(T 4 )

(7.51)

Replacing the expression of µ(T ) into U (T ), one gets : 5/2 π

U (T ) − U (T = 0) KεF

2

6

Å

kB T εF

ã2 = N εF

π2 4

Å

kB T εF

ã2 (7.52)

One deduces Cv : 3 dU π2 = N kB · Cv = dT 2 3

Å

kB T εF

ã (7.53)

Chapter 8

Elements of Bands Theory and Crystal Conductivity The previous chapter on free fermions allowed us to interpret some properties of metals. We considered that each electron was conﬁned within the solid by potential barriers, but did not introduce any speciﬁc description of its interactions with the lattice ions or with the other electrons. This model is not very realistic, in particular it cannot explain why some solids are conducting the electrical current and others are not : at room temperature, the conductivity of copper is almost 108 siemens (Ω−1 · m−1 ) whereas that of quartz is of 10−15 S, and of teﬂon 10−16 S. This physical parameter varies on 24 orders of magnitude according to the material, which is absolutely exceptional in nature. Thus we are going to reconsider the study of electrons in solids from the start and to ﬁrst address the characteristics of a solid and of its physical description. As we now know, the ﬁrst problem to solve is a quantum mechanical one, and we will rely on your previous courses. In particular we will be concerned by a periodic solid, that is, a crystal (§ 8.1). We will see that the electron eigenstates are regrouped into energy bands, separated by energy gaps (§ 8.2 and § 8.3). We will have to introduce here very simplifying hypotheses : in Solid State Physics courses, more general assumptions are used to describe electronic states of crystals. Using Statistical Physics (§ 8.4) we will see how electrons are ﬁlling energy bands. We will understand why some materials are conductors, other ones insulators. We will go more into details on the statistics of semiconductors,

175

176

Chapter 8. Elements of Bands Theory and Crystal Conductivity

of high importance in our everyday life, and will end by a brief description of the principles of a few devices based on semiconductors (§ 8.5).

8.1

What is a Solid, a Crystal ?

A solid is a dense system with a speciﬁc shape of its own. This deﬁnition expresses mechanical properties ; but one can also refer to the solid’s visual aspect, shiny in the case of metals, transparent in the case of glass, colored for painting pigments. These optical properties, as well as the electrical conduction and the magnetic properties, are due to the electrons. In a macroscopic solid with a volume in the cm3 range, the numbers of electrons and nuclei present are of the order of the Avogadro number N = 6.02 × 1023 . The hamiltonian describing this solid contains for each particle, either electron or nucleus, its kinetic energy, and its potential energy arising from its Coulombic interaction with all the other charged particles. It is easily guessed that the exact solution will be impossible to get without approximations (and still very diﬃcult to obtain using them !). One ﬁrst notices that nuclei are much heavier than electrons and, consequently, much less mobile : either the nuclei motion will be neglected by considering that their positions remain ﬁxed, or the small oscillations of these nuclei around their equilibrium positions (solid vibrations) will be treated seˆ will parately from the electron motions, so that the electron hamiltonian H ˆ will then be given by not include the nuclei kinetic energy. H ˆ = H

NZ N NZ N Z ˆ2 pi 1 1 Ze2 1 e2 − + n | 2 4πε0 2m 4πε0 i=1 n=1 |ri − R |r − rj | i=1 i,j=1 i

(8.1)

n and the where Z is the nuclear charge ; the nth nucleus is located at R electron i at ri . The second term expresses the Coulombian interaction of all the electrons with all the nuclei, the third one the repulsion among electrons. Besides, we will assume that the solid is a perfect crystal, that is, its atoms are n of a three-dimensional periodic lattice. This regularly located on the sites R will simplify the solution of the Quantum Mechanical problem, but many of our conclusions will also apply to disordered solids. ˆ depends of their probabilities of The repulsion term between electrons in H presence, in turn given by the solution of the hamiltonian : this is a diﬃcult self-consistent problem. The Hartree method, which averages the electronic repulsion, and the Hartree-Fock method, which takes into account the wave function antisymmetrization as required by the Pauli principle, are used to

The Eigenstates for the Chosen Model

177

solve this problem by iteration, but they lead to complicated calculations. Phenomena like ferromagnetism, superconductivity, Mott insulators, mixedvalence systems, are based on the interaction between electrons ; however in most cases one can account of the repulsion between electrons in a crystal through an average one-electron periodic term and we will limit ourselves to this situation in the present course. Then it will be possible to split the ˆ into one-electron terms, this is the independent-electron aphamiltonian H proximation : ˆ = H

N

ˆi h

(8.2)

i=1

with ˆ2 p ˆ hi = i + V(ri ) 2m

(8.3)

The potential V(ri ) has the lattice periodicity, it takes into account the attractions and repulsions on electron i and is globally attractive. The eigenvalues ˆ are the sums of the eigenenergies of the various electrons. of H ˆ i . We will Our task will consist in solving the one-electron hamiltonian h introduce an additional simpliﬁcation by assuming that the lattice is onedimensional, i.e., V(ri ) = V(xi ). This will allow us to extract the essential characteristics, at the cost of simpler calculations. In the “electron in the box” model that we have used up to now, V(xi ) only expressed the potential barriers limiting the solid, and thus leveled out the potential of each atom. Here we will introduce the attractive potential on each lattice site. The energy levels will derive from atomic levels, rather than from the free electron states which were the solutions in the “box” case.

8.2 8.2.1

The Eigenstates for the Chosen Model Recall : the Double Potential Well

First consider an electron on an isolated atom, located at the origin. Its hamiltonian ˆ h is given by 2 ˆ = pˆ + V (x) h 2m

where V is the even Coulombian potential of the nucleus (Fig. 8.1).

(8.4)

178

Chapter 8. Elements of Bands Theory and Crystal Conductivity

One of its eigenstate is such that : ˆ hϕ(x) = ε0 ϕ(x)

(8.5)

ϕ(x) is an atomic orbital, of s-type for example, which decreases from the origin with a characteristic distance λ. We will assume that the other orbitals do not have to be considered in this problem, since they are very distant in energy from ε0 . V

λ

|ϕ|2

x

Fig. 8.1: Coulombian potential of an isolated ion and probability of presence of an electron on this ion. Let us add another nucleus, that we locate at x = d (Fig. 8.2). The hamiltonian, still for a single electron now sharing its presence between the two ions, 2 ˆ = pˆ + V (x) + V (x − d) h 2m

(8.6)

In the case of two 1s-orbital initial states, this corresponds to the description of the H+ 2 ion (two protons, one electron). One follows the standard method on the double potential well of the Quantum Mechanics courses and deﬁnes |ϕL and |ϕR , the atomic eigenstates respectively located on the left nucleus (at x = 0 ) and the right nucleus (at x = d) : ⎧ñ ô ⎪ pˆ2 ⎪ ⎪ + V (x) |ϕL = ε0 |ϕL ⎪ ⎪ ⎨ 2m (8.7) ⎪ ï 2 ò ⎪ ⎪ pˆ ⎪ ⎪ + V (x − d) |ϕR = ε0 |ϕR ⎩ 2m

The Eigenstates for the Chosen Model

179

V

d

0

x ε0 + A ε0 − A

Fig. 8.2: Potential due to two ions of the same nature, distant of d and corresponding bound states. Owing to the symmetry of the potential in (8.6) with respect to the point ˆ can be chosen either symmetrical x = d/2, whatever d the eigenfunctions of h or antisymmetrical with respect to this mid-point. However, if d is very large ˆ has two with respect to λ, the coupling between the two sites is very weak, h eigenvalues practically equal to ε0 and if at the initial time the electron is on one of the nuclei, for example the left one L, it will need a time overcoming the observation possibilities to move to the right one R. One can then consider that the atomic states (8.7) are the eigenstates of the problem (although strictly they are not). On the other hand, you learnt in Quantum Mechanics that if d is not too large with respect to λ, the L and R wells are coupled by the tunnel eﬀect, which is expressed through the coupling matrix element ˆ L , −A = ϕR |H|ϕ

A>0

(8.8)

which decreases versus d approximately like exp(−d/λ). ˆ are equal The energy degeneracy is then lifted : the two new eigenvalues of h 1 to ε0 − A, corresponding to the symmetrical wave function |ψS = √ (|ϕL + 2 1 |ϕR ), and ε0 + A, of antisymmetrical wave function |ψA = √ (|ϕL − |ϕR ) 2 (Fig. 8.2). These results are valid for a weak enough coupling (A |ε0 |). For a stronger coupling (d ∼ λ) one should account for the overlap of the two atomic functions, expressed by ϕL |ϕR diﬀerent from zero. Anyway the eigenstates are delocalized between both wells. For H+ 2 they correspond to the bonding and antibonding states of the chemical bound.

180

8.2.2

Chapter 8. Elements of Bands Theory and Crystal Conductivity

Electron on an Inﬁnite and Periodic Chain

In the independent electrons model, the hamiltonian describing a single electron in the crystal is given by +∞ 2 ˆ crystal = pˆ + V (x − xn ) h 2m n=−∞

(8.9)

where xn = nd, is the location of the nth nucleus. The potentials sum is periodic and plays the role of V(ri ) in formula (8.3) (Fig. 8.3). V

0

d

2d

nd x

Fig. 8.3 : Potential periodic of a crystal. One is looking for the stationary solution, such that ˆ crystal ψ(x) = εψ(x) h

(8.10)

If the nuclei are far enough apart, there is no coupling : although the eigenstates are delocalized, if at the initial time the electron is on a given nucleus, an inﬁnite time would be required for this electron to move to the neighboring nucleus. The state of energy ε = ε0 is N times degenerate. On the other hand, in the presence of an even weak coupling, the eigenstates are delocalized. We are going to assume that the coupling is weak and will note : ˆn + ˆ crystal = h V (x − n d) (8.11) h n =n

to show the atomic hamiltonian on the speciﬁc site n. To solve the eigenvalue problem, by analogy with the case of the double potential well, we will work in a basis of localized states, which are not the eigenstates of the considered inﬁnite periodic chain. Now every site plays a similar role, the problem is invariant under the translations which change x into x + nd = x + xn . The general basis state is written |n, it corresponds to the wave function φn (x) located around xn and is deduced through the translation xn from the function centered at the origin, i.e., φn (x) = φ0 (x − xn ). These states are built through orthogonalization from the atomic functions ϕ(x − xn ) in order to constitute an orthonormal basis : n |n = δnn (this is

The Eigenstates for the Chosen Model

181

the so-called “Hückel model”). The functions φn (x) depend on the coupling. When the atoms are very distant, φn (x) is very close to ϕ(x − xn ), the atomic wave function. In a weak coupling situation, where the tunnel eﬀect only takes place between ﬁrst neighbors, φn (x) is still not much diﬀerent from the atomic function. One looks for a stationary solution of the form |ψ = cn |n

(8.12)

n

One needs to ﬁnd the coeﬃcients cn , expected to all have the same norm, since every site plays the same role : ˆ hcrist

+∞

+∞ cn |n = ε cn |n

n=−∞

(8.13)

n=−∞

One proceeds by analogy with the solution of the double-well potential problem : one assumes that the hamiltonian only couples the ﬁrst neighboring sites. Then the coupling term is given by ˆ crist |n + 1 = n − 1|h ˆ crist |n = −A n|h

(8.14)

For atomic s-type functions ϕ(x) A is negative. By multiplying (8.13) by the bra n|, one obtains ˆ cryst |n − 1 + cn n|h ˆ cryst |n + cn+1 n|h ˆ cryst |n + 1 = cn ε cn−1 n|h Now ˆ cryst |n = n|h ˆ n |n + n| n|h

(8.15)

V (x − n d)|n

n =n

(8.16)

= ε0 + α ε0 The term α, which is the integral of the product of |φn (x)|2 by the sum of potentials of sites distinct from xn , is very small in weak coupling conditions [it varies in exp(−2d/λ), whereas A is in exp(−d/λ)]. It will be neglected with respect to ε0 . The coeﬃcients cn , cn−1 and cn+1 are then related through −cn−1 A + cn ε0 − cn+1 A = cn ε

(8.17)

One obtains analogous equations by multiplying (8.13) by the other bras, whence ﬁnally the set of coupled equations : ⎧ ⎪ ⎨. . . . . . (8.18) −cn−1 A + cn ε0 − cn+1 A = cn ε ⎪ ⎩ − cn A + cn+1 ε0 − cn+2 A = cn+1 ε

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8.2.3

Energy Bands and Bloch Functions

The system (8.18) assumes a nonzero solution if one chooses cn as a phase : cn = exp(ik · nd)

(8.19)

where k, homogeneous to a reciprocal length, is a wave vector and nd the coordinate of the considered site. By substituting (8.19) into (8.17) one obtains the dispersion relation, or dispersion law, linking ε and k : ε(k) = ε0 − 2A cos kd

(8.20)

ε(k) is an even function, periodic in k, which describes the totality of its values, i.e. the interval [ε0 − 2A, ε0 + 2A], when k varies from −π/d to π/d (Fig. 8.4). The domain of accessible energies ε constitutes an allowed band ; a value of ε chosen outside of the interval [ε0 − 2A, ε0 + 2A] does not correspond to a kvalue, it belongs to a forbidden domain. The range −π/d ≤ k < π/d is called “the ﬁrst Brillouin zone,” a concept introduced in 1930 by Léon Brillouin, a French physicist ; it is suﬃcient to study the physics of the problem. Note that ε(k)

−π/d

π/d

k

ε0 + 2 A ε0

4A

ε0 − 2 A

Fig. 8.4 : Dispersion relation in the crystal. in the neighborhood of k = 0, ε(k) ε0 − 2A + Ad2 k 2 + . . . which looks like the free electron dispersion relation, if an eﬀective mass m∗ is deﬁned such that 2 /2m∗ = Ad2 . This mass also expresses the smaller or larger diﬃculty for an electron to move under the application of an external electric ﬁeld, due to the interactions inside the solid it is subjected to. The wave function associated with ε(k) is given by ψk (x) =

+∞ n=−∞

exp iknd · φ0 (x − nd)

(8.21)

The Eigenstates for the Chosen Model

183

This expression can be transformed into ψk (x) = exp ikx ·

+∞

exp −ik(x − nd) · φ0 (x − nd)

(8.22)

n=−∞

which is the product of a plane wave by a periodic function uk (x), of period d : indeed, for n integer, uk (x + n d) =

+∞

exp −ik(x + n d − nd) · φ0 (x + n d − nd)

n=−∞

=

+∞

(8.23)

exp −ik(x − n d) · φ0 (x − n d) = uk (x)

n =−∞

This type of solution of the Schroedinger stationary equation ψk (x) = eikx · uk (x)

(8.24)

can be generalized, for a three-dimensional periodic potential V (r), into

ψk (r) = eik·r · uk

(8.25)

This is the so-called Bloch function, from the name of Felix Bloch, who, in his doctoral thesis prepared in Leipzig under the supervision of Werner Heisenberg and defended in July 1928, was the ﬁrst to obtain the expression of a crystal wave function, using a method analogous to the one followed here. The expression results from the particular form of the diﬀerential equation to be solved (second derivative and periodic potential) and is justiﬁed using the Floquet theorem. Expression (8.21) of ψk (x), established in the case of weak coupling between neighboring atomic sites, can be interpreted as a Linear Combination of Atomic Orbitals (LCAO), since the φ0 ’s then do not diﬀer much from the atomic solutions ϕ. This model is also called “tight binding” as it privileges the atomic states. When the atomic states of two neighboring sites are taken as orthogonal, this is the so-called “Hückel model.” The ψk (x) functions are delocalized all over the crystal. The extreme energy states, at the edges of the energy band, correspond to particular Bloch functions : the minimum energy state ε0 − 2A is associated with ψ0 (x) = +∞ n=−∞ φ0 (x−nd), the addition in phase of each localized wave function : this is a situation comparable to the bonding symmetricalstate of the H+ 2 ion ; the n state of energy ε0 + 2A is associated with ψπ/d (x) = +∞ n=−∞ (−1) φ0 (x− nd), alternate addition of each localized wave function and is analogous to the antibonding state of the H+ 2 ion.

184

Chapter 8. Elements of Bands Theory and Crystal Conductivity

One can notice that the band width 4A obtained for an inﬁnite chain is only twice the value of the energy splitting for a pair of coupled sites. The reason is that, in the present model, a given site only interacts with its ﬁrst neighbors and not with the more distant sites Note : the dispersion relation (8.20) and the wave function (8.24), that we obtained under very speciﬁc hypotheses, can be generalized in the case of three-dimensional periodic solids, as can be found in basic Solid State Physics courses, like the one taught at Ecole Polytechnique.

8.3 8.3.1

The Electron States in a Crystal Wave Packet of Bloch Waves

Bloch waves of the (8.21) or (8.24) type are delocalized and thus cannot describe the amplitude of probability of presence for an electron, as these waves cannot be normalized : although a localized function φ0 (x) can be normed, ψk (x) spreads to inﬁnity. To describe an electron behavior, one has to build a packet of Bloch waves

+∞

Ψ(x, t) = −∞

g(k)eikx uk (x)e−iε(k)t/ dk

(8.26)

centered around a wave vector k0 , of extent ∆k small with respect to π/d. From the Heisenberg uncertainty principle, the extent in x of the packet will be such that ∆x · ∆k ≥ 1, i.e., ∆x d : the Bloch wave packet is spreading over a large number of elementary cells.

8.3.2

Resistance ; Mean Free Path

It can be shown that an electron described by a wave packet (8.26), in motion in a perfect inﬁnite crystal, keeps its velocity value for ever and thus is not diﬀused by the periodic lattice ; in fact the Bloch wave solution already includes all these coherent diﬀusions. Then there is absolutely no resistance to the electron motion or, in other words, the conductivity is inﬁnite. modiﬁes the electron energy levels ; the proA static applied electric ﬁeld E blem is complex, one should also account for the eﬀect of the lattice ions and interact with the electron. which are also subjected to the action of E Experimentally one notices in a macroscopic crystal that the less its defects,

The Electron States in a Crystal

185

like dislocations or impurities, the higher its conductivity. An inﬁnite perfect three-dimensional crystal would have an inﬁnite conductivity, a result predicted by Felix Bloch as early as 1928. (These simple considerations do not apply to the recently elaborated, low-dimensional, systems.) The real ﬁnite conductivity of crystals arises from the discrepancies to the perfect inﬁnite lattice : thermal motion of the ions, defects, existence of the surface. In fact in copper at low temperature one deduces from the conductivity value σ = ne2 τ /m the mean time τ between two collisions and, introducing the Fermi velocity vF (see §7.1.1), one obtains a mean free path = vF τ of the order of 40 nm, that is, about 100 interatomic distances. This does show that the lattice ions are not responsible for the diﬀusion of conduction electrons.

8.3.3

Finite Chain, Density of States, Eﬀective Mass

V (x) 0

d

2d

Nd

(N + 1)d x

Fig. 8.5 : Potential for a N -ions chain. A real crystal is spreading over a macroscopic distance L = N d (Fig. 8.5) (the distance L is measured with an uncertainty much larger than the interatomic distance d ; one can thus always assume that L contains an integer number N of interatomic distances). Obviously, if only the coupling to the nearest neighbors is considered, for most sites there is no change with respect to the case of the inﬁnite crystal. Thus the eigenstates of the hamiltonian for the ﬁnite crystal : 2 ˆ crystal = pˆ + V (x − xn ) h 2m n=1 N

(8.27)

should not diﬀer much from those of the hamiltonian (8.9) for the inﬁnite crystal : we are going to test the Bloch function obtained for the inﬁnite crystal as a solution for the ﬁnite case. On the other hand, for a macroscopic dimension L the exact limit conditions do not matter much. It is possible, as in the case of the free electron model, to close the crystal on itself by superposing the sites 0 and N d and to express

186

Chapter 8. Elements of Bands Theory and Crystal Conductivity

the periodic limit conditions of Born-Von Kármán : ψ(x + L) = ψ(x) ψk (x + L) = e

ik(x+L)

(8.28) uk (x + L) = e

ikL

ψk (x)

(8.29)

as L contains an integer number of atoms, i.e. an integer number of atomic periods d, and uk (x) is periodic. One thus ﬁnds the same quantization condition as for a one-dimensional free particle : k = p2π/L, where p is an integer, associated with an one-dimension orbital density of states L/2π ; the corresponding three-dimensional density of states is Ω/(2π)3 , where Ω is the crystal volume. These densities are doubled when one accounts for the electron spin. The solutions of (8.27) are then ⎧ kd ⎪ ⎨ε = ε0 − 2A cos N 1 ψk (x) = √N n=1 exp iknd · φ0 (x − nd) ⎪ ⎩ with k = p · 2π p integer L

(8.30)

When comparing to the results (8.20) and (8.21) for the inﬁnite crystal, one ε is now quanﬁrst notices that ψk (x) as given by (8.30) is normed. Moreover π π tized (Fig. 8.6) : to describe the ﬁrst Brillouin zone − ≤ k < all the d d N N diﬀerent states are obtained for p ranging between − and + − 1 (or 2 2 N −1 N −1 between − and − 1 if N is odd, but because of the order of 2 2 magnitude of N , ratio of the macroscopic distance L to the atomic one d, this does not matter much !). On this interval there are thus N possible states, i.e., as many as the number of sites. Half of the states correspond to waves propagating toward increasing x’s (k > 0), the other ones to waves propagating toward decreasing x’s. Note that if the states are uniformly spread in k, their distribution is not uniform in energy. One deﬁnes the energy density of states D(ε) such that the number of accessible states dn, of energies between ε and ε + dε, is equal to dn = D(ε)dε

(8.31)

One will note that, at one dimension, to a given value of ε correspond two opposite values of kx , associated with distinct quantum states. By observing Fig. 8.6, it can be seen that D(ε) takes very large values at the vicinity of the band edges (ε close to ε0 ± 2A) ; in fact at one dimension D(ε) has a speciﬁc behavior [D(ε) diverges at the band edges, but the integral ε D(ε )dε (8.32) εmin

The Electron States in a Crystal

187

ε(k) ε +2A

ε0 − 2 A −π/d

2π 2π = Nd L

0

π/d

k

Fig. 8.6 : Allowed states in the energy band, for a N electrons system.

remains ﬁnite]. When this model is extended to a three-dimensional system, one ﬁnds again a “quasicontinuous” set of very close accessible states, uniformly located in k (but not in energy), in number equal to that of the initial orbitals (it was already the case for two atoms for which we obtained the bonding and antibonding states). V x 2s

1s

2s band

1s band

Fig. 8.7: In a N atom crystal, to each atomic state corresponds a band that can accommodate 2N times more electrons than the initial state. In a more realistic case, the considered atom has several energy levels (1s, 2s, 2p, ...). Each energy level gives rise to a band, wider if the atomic state lies higher in energy (Fig. 8.7). The widths of the bands originating from the deep atomic states are extremely small, one can consider that these atomic levels do not couple. On the other hand, the width of the band containing the electrons involved in the chemical bond is always of the order of several electron-volts.

188

Chapter 8. Elements of Bands Theory and Crystal Conductivity

Again each band contains N times more orbital states than the initial atomic state, that is, 2N times more when the electron spin is taken into account. The density of states in k is uniform. One deﬁnes an energy density of states D(ε) such that Å dn = D(ε)dε = 2

L 2π

ã3

d3k at three dimensions

(8.33)

The area under the curve representing the density of states of each band is equal to 2N times the number of electrons acceptable in the initial atomic orbital : 2N for s states, 6N for p states, etc. (Fig. 8.8). The allowed bands are generally separated by forbidden bands, except when there is an overlap between the higher energy bands, which are broadened through the coupling (3s and 3p on the Fig. 8.8). 1s D(ε)

2p 2s 3s, 3p

ε Fig. 8.8 : Density of states of a crystal. It can be shown that in the vicinity of a band extremum, the tangent of the D(ε) curve is vertical. In particular, at the bottom of a band the density of state is rather similar to the one of a free electron in a box [see (6.64)], for √ Ω which we obtained D(ε) = 2 3 (2m)3/2 ε, m being the free electron mass. 2π Here again one will deﬁne the eﬀective mass in a given band, this time from the curvature of D(ε) [this mass is the same as the one in § 8.2.3, introduced from the dispersion relation (8.20)]. Indeed, setting D(ε)

√ Ω (2m∗ )3/2 ε − ε0 2π 2 3

(8.34)

in the vicinity of ε0 at the bottom of the band is equivalent to taking for dispersion relation ε − ε0

2 k 2 2m∗

,

i.e.,

2 d2 ε = m∗ dk 2

(8.35)

Statistical Physics of Solids

189

The eﬀective mass is linked to the band width : from the dispersion relation (8.20) for instance, one obtains around k = 0 2Ad2 1 = ∗ m 2

(8.36)

One can understand that the electrons of the deep bands are “very heavy,” and thus not really mobile inside the lattice, that is, they are practically localized. Note : in non crystalline (or amorphous) solids, like amorphous silicon, the coupling between the atomic states also produces energy bands.

8.4

Statistical Physics of Solids

Once the one-electron quantum states determined, one has to locate the N electrons in these states, in the framework of the independent electrons model.

8.4.1

Filling of the Levels

At zero temperature, the Pauli principle dictates one to ﬁll the levels from the lowest energy one, putting two electrons, with diﬀerent spin states, on each orbital state until the N electrons are exhausted. We now take two examples, that of lithium (Z = 3), and that of carbon (Z = 6) crystallized into the diamond form. In lithium, each atom brings 3 electrons, for the whole solid there are 3N electrons to be located. The 1s band is ﬁlled by 2N electrons. The complementary N electrons are ﬁlling the 2s band up to its half, until the energy ε = εF , the Fermi energy, which lies 4.7 eV above the bottom of this band. Above εF , any state is empty at T = 0 K. When a small amount of energy is brought to the solid, for example using a voltage source of a few volts, the electrons in the 2s band in the vicinity of εF can occupy states immediately above εF , whereas 1s electrons cannot react, having no empty states in their neighborhood. This energy modiﬁcation of electrons in the 2s band produces a global macroscopic reaction, expressed by an electric current, to the electric ﬁeld : lithium is a metal or a conductor, the 2s band is its conduction band. In diamond, the initial electronic states are not the 2s and 2p states of the atomic carbon but their sp3 tetrahedral hybrids (as in CH4 ) which separate into bonding hybrids with 4 electrons per atom (including spin) and antibon-

190

Chapter 8. Elements of Bands Theory and Crystal Conductivity

ding hybrids also with 4 electrons. In the crystal, the bonding states constitute a 4N -electron band, the same is true for the antibonding states. The 4N electrons of the N atoms exactly ﬁll the bonding band, which contains the electrons participating in the chemical bond, whence the name of valence band. The band immediately above, of antibonding type, is the conduction band. It is separated from the valence band by the band gap energy, noted Eg : Eg = 5.4 eV for diamond and at T = 0 K the conduction band is empty, the valence band having exactly accommodated all the available electrons. Under the eﬀect of an applied voltage of a few volts, the valence electrons cannot be excited through the forbidden band, there is no macroscopic reaction, the diamond is an insulator. Like carbon, silicon and germanium also belong to column IV of the periodic table, they have the same crystalline structure, they are also insulators. However these latter atoms are heavier, their crystal unit cells are larger and the characteristic energies smaller : thus Eg = 1.1 eV in silicon and 0.75 eV in germanium. Remember that a full band has no conduction, the application of an electric ﬁeld producing no change of the macroscopic occupation of the levels. This is equally valid for the 1s band of the lithium metal and at zero temperature for the valence band of insulators like Si or Ge.

8.4.2

Variation of Metal Resistance versus Temperature

We now are able to justify the methods and results on metals of the previous chapter, by reference to the case of lithium discussed above : their bands consist in one or several full bands, inactive under the application of an electric ﬁeld, and a conduction band partly ﬁlled at zero temperature. The density of states at the bottom of this latter band is analogous to that of an electron in a box, under the condition that the free electron mass is replaced by the eﬀective mass at the bottom of the conduction band. Indeed we already anticipated this result when we analyzed the speciﬁc heats of copper and silver (7.2.2). The electrical conductivity σ is related to the linear eﬀect of an applied ﬁeld E on the conduction electrons. An elementary argument allows one to establish its expression σ = ne2 τ /m∗ , where n is the electron concentration in the conduction band, e the electron charge, τ the mean time between two collisions and m∗ the eﬀective mass. A more rigorous analysis, in the framework of the transport theory, justiﬁes this formula which is far from being obvious : the only excitable electrons are in the vicinity of the Fermi level, their velocity in the absence of electric ﬁeld is vF . The phenomenological τ term takes into account the various mechanisms

Statistical Physics of Solids

191

which limit the conductivity : in fact the probabilities of these diﬀerent independent mechanisms are additive, so that 1 1 1 = + + ··· τ τvibr τimp

(8.37)

1/τvibr corresponds to the probability of diﬀusion by the lattice vibrations, which increases almost linearly with the temperature T , 1/τimp to the diffusion by neutral or ionized impurities, a mechanism present even at zero temperature. Then 1 m∗ =ρ= 2 σ ne

Å

1 τvibr

+

1 τimp

ã + ···

(8.38)

ρ = ρvibr (T ) + ρimp + · · · Thus the resistivity terms arising from the various mechanisms add up, this is the empirical Matthiessen law, stated in the middle of the 19th century. Consequently, the resistance of a metal increases with temperature according to a law practically linear in T , the dominating mechanism being the electron’s diﬀusion on the lattice vibrations. This justiﬁes the use of the platinum resistance thermometer as a secondary temperature standard between −183◦ C and +630◦ C.

8.4.3

Variation of Insulator’s Conductivity Versus Temperature ; Semiconductors

a) Concentrations of Mobile Carriers in Equilibrium at Temperature T We have just seen that at zero temperature an electric ﬁeld has no eﬀect on an insulator, since its valence band is totally full and its conduction band empty. However there are processes to excite such systems. The two main ones are : – an optical excitation : a photon of energy hν > Eg will promote an electron from the valence band into the conduction band. This is the absorption phenomenon, partly responsible for the color of these solids. It creates a nonequilibrium thermal situation. – a thermal excitation : the thermal motion is also at the origin of transitions from the valence band to the conduction band. We are now going to detail this latter phenomenon, which produces a thermal equilibrium at temperature T .

192

Chapter 8. Elements of Bands Theory and Crystal Conductivity

When an electron is excited from the valence band, this band then makes a transition from the state (full band) to the state (full band minus one electron), which allows one to deﬁne a quasi-particle, the “hole,” which is the lack of an electron. We already noted in chapter 7.1.1 the symmetry, due to the shape of the Fermi distribution, between the occupation factor f (ε) of energy states above µ, and the probability 1 − f (ε) of the states below µ for being empty. The following calculation will show the similarity between the electron or hole role. The electrons inside the insulator follow the Fermi-Dirac statistics and at zero temperature the conduction band is empty and the valence band full. This means that εF lies between εv , the top of the valence band and εc , the minimum of the conduction band. At weak temperature T (kB T Eg ) one expects that the chemical potential µ still lies inside the energy band gap. We are going to estimate n(T ), the number of electrons present at T in the conduction band and p(T ), the number of electrons missing (or holes present) in the valence band, and to deduce the energy location µ(T ) of the chemical potential. These very few carriers will be able to react to an applied electric : indeed in the conduction band the electrons are very scarce at usual ﬁeld E temperatures and they are able to ﬁnd vacant states in their close neighbo ; in the same way the presence of several holes in rhood when excited by E the valence band permits the valence electrons to react to E. The volume concentration n(T ) of electrons present in the conduction band at temperature T in the solid of volume Ω is such that εcmax εcmax 1 dε (8.39) D(ε)f (ε)dε = D(ε) Ωn(T ) = exp β(ε − µ) + 1 εc εc where εc is the energy of the conduction band minimum, εcmax the energy of its maximum. One assumes, and it will be veriﬁed at the end of the calculation, that µ is distant of this conduction band of several kB T , which leads to β(εc − µ) 1, so that one can approximate f (ε) by a Boltzmann factor. Then Ωn(T )

εcmax

D(ε)e−β(ε−µ) dε

(8.40)

εc

Because of the very fast exponential variation, the integrand has a signiﬁcant value only if ε − εc does not exceeds a few kB T . Consequently, one can extend the upper limit of the integral to +∞, as the added contributions are absolutely negligible. Moreover, only the expression of D(ε) at the bottom of the conduction band really matters : we saw in (8.34) that an eﬀective mass, here mc , can be introduced, which expresses the band curvature in the vicinity of

Statistical Physics of Solids

193

ε = εc . Whence

∞

√ 1 (2mc )3/2 ε − εc e−β(ε−µ) dε 2 3 2π εc ∞ √ −x 1 −β(εc −µ) 3/2 e (2mc kT ) xe dx = 2π 2 3 0 ã ã Å Å 2πmc kB T 3/2 εc − µ =2 exp − h2 kB T

n(T )

(8.41)

(8.42)

In the same way one calculates the hole concentration p(T ) in the valence band, the absence of electrons at temperature T being proportional to the factor 1 − f (ε) : εv Ωp(T ) = D(ε)[1 − f (ε)]dε (8.43) εvmin

One assumes that µ is at a distance of several kB T from εv , the maximum of the valence band, of minimum εvmin . The neighborhood of εv is the only region that will intervene in the integral calculation because of the fast decrease of 1 − f (ε) e−β(µ−ε) . In this region D(ε) is parabolic and is approximated by D(ε)

√ Ω (2mv )3/2 εv − ε 2 3 2π

(8.44)

where mv > 0, the “hole eﬀective mass,” expresses the curvature of D(ε) in the neighborhood of the valence band maximum. One gets : Å

2πmv kB T p(T ) = 2 h2

ã3/2

ã Å µ − εv exp − kB T

(8.45)

In the calculations between formulae (8.40) and (8.45) we did not make any assumption about the way n(T ) and p(T ) are produced. These expressions are general, as is the product of (8.42) by (8.45) : Å n(T ) · p(T ) = 4

2πkB T h2

ã3

ã Å Eg (mc mv )3/2 exp − kB T

(8.46)

with Eg = εc − εv . b) Pure (or Intrinsic) Semiconductor In the case of a thermal excitation, for each electron excited into the conduction band a hole is left in the valence band. One deduces from (8.46) : Å n(T ) = p(T ) = ni (T ) = 2

2πkB T h2

ã3/2

ã Å Eg (mc mv )3/4 exp − 2kB T

(8.47)

194

Chapter 8. Elements of Bands Theory and Crystal Conductivity

For this intrinsic mechanism, the variation of ni (T ) versus temperature is dominated by the exp(−Eg /2kB T ) term. In fact, here the thermal mechanism creates two charge carriers with a threshold energy Eg . The very fast variation of the exponential is responsible for the distinction between insulators and semiconductors. Indeed at room temperature T = 300 1 eV, let us compare the diamond K, which corresponds to approximately 40 situation where Eg = 5.2 eV [exp(−Eg /2kB T ) = exp(−104) = 6.8 × 10−46 ] and that of silicon, where Eg = 1.1 eV [exp(−22) = 2.8 × 10−10 ]. In 1 cm3 of diamond, there is no electron in the conduction band, whereas ni (T ) is equal to 1.6 × 1016 m−3 at 300 K in silicon. A semiconductor is an insulator with a “not too wide” band gap, which allows the excitation of some electrons into the conduction band at room temperature. The most commonly used semiconductors, apart from silicon, are germanium and gallium arsenide GaAs (Eg = 1.52 eV). The chemical potential of an intrinsic semiconductor is determined by substituting expression (8.47) of ni (T ) into (8.41). One obtains : µ=

3 mv εc + εv + kB T ln 2 4 mc

(8.48)

The Fermi level εF , which is the chemical potential at T = 0 K, lies in the middle of the band gap and µ does not move much apart from this location when T increases (kB T Eg ). Consequently, the hypotheses that µ lies very far from the band edges as compared to kB T , made when calculating n(T ) and p(T ), are valid. Under the eﬀect of an external electric ﬁeld, both electrons and holes participate in the electrical current : indeed the current includes contributions from both the conduction and the valence bands. As will be studied in Solid State Physics courses, the holes behave like positive charges and the total electrical conductivity in the semiconductor is the sum of the valence band and conduction band currents : σ = σn + σp ã Å τn τp = ne2 + pe2 mc mv

(8.49)

Here τn (respectively τp ) is the electron (respectively hole) collision time. The temperature variation of σ in an intrinsic semiconductor is dominated by the variation of exp(−Eg /2kT ), the other factors varying like powers of T . Consequently, σ increases very fast with T , thus its reciprocal, the resistivity and the measured resistance of a semiconductor decrease with temperature because the number of carriers increases.

Statistical Physics of Solids

195

c) Doped Semiconductor The density of intrinsic carriers at room temperature is very weak and practically impossible to attain : for instance in silicon about 1 electron out of 1012 is in the conduction band due to thermal excitation. To achieve the intrinsic concentration the studied sample should contain neither impurities nor defects ! It is more realistic to control the conductivity by selected impurities, the dopants. Thus phosphorus atoms can be introduced into a silicon crystal : phosphorus is an element of column V of the periodic table, which substitutes for silicon atoms in some lattice sites. Each phosphorus atom brings one more electron than required by the tetrahedral bonds of Si. At low temperature this electron remains localized near the phosphorus nucleus, on a level located 44 meV below the bottom of the conduction band. When the temperature increases, the electron is ionized into the conduction band. Phosphorus is a donor. The concentration of conduction electrons will be determined by the concentration ND in P atoms, of typical value ND ∼ 1022 m−3 , chosen much larger than the intrinsic and the residual impurity concentrations, but much smaller than the Si atoms concentration. In such a material, because of (8.46), the hole concentration is very small with respect to ni , the current is carried by the electrons, the material is called of n-type. In a symmetrical way, the substitution in the silicon lattice of a Si atom by a boron atom creates an acceptor state : boron, an element of column III of the periodic table, only has three peripheral electrons and thus must capture one extra electron in the valence band to satisfy the four tetrahedral bonds of the crystal. This is equivalent to creating a hole, which, at low temperature, is localized 46 meV above the silicon valence band and at higher temperature is delocalized inside the valence band. In boron-doped silicon, the conductivity arises from the holes, the semiconductor is of p-type. In Solid State Physics courses, the statistics of doped semiconductors is studied in more detail : one ﬁnds in particular that the chemical potential µ moves versus temperature inside the band gap (in a n-type material, at low temperature it is close εc , in the vicinity of the donors levels and returns to the middle of the band gap, like in an intrinsic material, at very high temperature). The semiconductors constituting the electronic compounds are generally doped.

196

8.5

Chapter 8. Elements of Bands Theory and Crystal Conductivity

Examples of Semiconductor Devices

Here we only give a ﬂavor of the variety of conﬁgurations and uses. In specialized Solid State Physics courses, one will describe in more details the semiconductors properties. The aim here is to show that, just using the basic elements of the present chapter, one can understand the principle of operation of three widespread devices.

8.5.1

The Photocopier : Photoconductivity Properties

Illuminating a semiconductor by light of energy larger than Eg promotes an electron into the conduction band and leaves a hole in the valence band. The conductivity increase is ∆σ, proportional to the light ﬂux. In the xerography process, the photocopier drum is made of selenium which, under the illumination eﬀect, shifts from a very insulator state to a conducting state. This property is utilized to ﬁx the pigments (the toner black powder) which are next transferred to the paper and then cooked in an oven to be stabilized on the paper sheet.

8.5.2

The Solar Cell : an Illuminated p − n Junction

When, in the same solid, a p region is placed next to a n region, the electrons of the n region diﬀuse toward the holes of the p region : some of these mobile charges neutralize and a so-called space charge region becomes emptied of mobile carriers. The dopants still remain in the region, their charged nuclei create an electric ﬁeld which prevents any additional diﬀusion. This eﬀect can also be interpreted from the requirement of equal chemical potentials in the n and p region at equilibrium : in the neighborhood (a few µm) of the i . If region of doping change, there prevails an intense internal electric ﬁeld E the sun is illuminating this region, the excited electron and hole, of opposite i . An external use circuit will take advantage of charges, are separated by E the current produced by this light excitation. We will reconsider these devices, called “solar cells,” at the end of next chapter when we will deal with thermal radiation.

Examples of Semiconductor Devices

8.5.3

197

Compact Disk Readers : Semiconductor Quantum Wells

It was seen in § 8.3.3 that starting from the energy levels of the constituting atoms one gets the energy bands of a solid. Their location depends of the initial atomic levels, thus of the chemical nature of the material. For almost thirty years it has been possible, in particular using Molecular Beam Epitaxy (MBE), to sequentially crystallize diﬀerent semiconductors, which have the same crystal lattice but diﬀerent chemical composition : a common example is GaAs and Ga1−x Alx As. In such semiconductor sandwiches the bottom of the conduction band (or the top of the valence band ) does not lie at the same energy in GaAs and in Ga1−x Alx As. One thus tailors quantum wells, to fashion. Their width usually ranges from 5 to 20 nm, their energy depth is of a few hundreds of meV, and the energies of their localized levels is determined by their “geometry,” as taught in Quantum Mechanics courses. These quantum wells are much used in optoelectronics, for example as diodes or lasers emitters, in the telecommunication networks by optical ﬁbers, as readers of compact disks, of code-bars in supermarkets, etc.

Summary of Chapter 8 One has ﬁrst to study the quantum states of a periodic solid. By analogy with the problem of the double potential well of the Quantum Mechanics course, one considers an electron on an inﬁnite periodic chain, in a model of independent electrons, of the Linear Combination of Atomic Orbitals type. The hamiltonian eigenstates are Bloch states, the eigenenergies are regrouped into allowed bands separated by forbidden bands (or band gaps). The important parameter is the coupling term between neighboring sites. In a perfect crystal, the velocity of an electron would be maintained forever. The ﬁnite macroscopic dimension of a crystal allows one to deﬁne a density of states and to show that the number of orbital states is the product of the number of cells in the crystal by the number of orbital states corresponding to each initial atomic state. Filling these states together with considering the spin properties, allows one to distinguish between insulators and metals. We calculated the number of mobile carriers of an insulator versus temperature. The resistance of an insulator exponentially decreases when T increases, the resistance of a metal increases almost linearly with T . A semiconductor is an insulator with a “rather small” band gap energy.

199

Chapter 9

Bosons : Helium 4, Photons, Thermal Radiation We learnt in chapter 5 that bosons are particles of integer or null spin, with a symmetrical wave function, i.e. which is unmodiﬁed in the exchange of two particles. Bosons can be found in arbitrary number on a one-particle state of given energy εk . We are going now to discuss in more detail the properties of systems following the Bose-Einstein statistics. We will begin with the case of material particles and take the example of 42 He, the helium isotope made up of two protons and two neutrons. The Bose-Einstein condensation occurring in this case is related to the helium property of superﬂuidity. Then we will study the statistics in equilibrium of photons, relativistic particles of zero mass. The photon statistics has an essential importance for our life on the earth through the greenhouse eﬀect. Moreover, it played a major conceptual part : indeed it is in order to interpret the experimental thermal radiation behavior that Max Planck (1858-1947) produced the quanta hypothesis (1901). Finally we will take a macroscopic point of view and stress the importance of thermal radiation phenomena in our daily experience.

9.1

Material Particles

The ability of bosons of given non-zero mass to gather on the same level under a low enough temperature yields to a phase transition, the Bose-Einstein condensation.

201

202

9.1.1

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Thermodynamics of the Boson Gas

Let us recall the average occupation number nk [Eq.(6.26)] by bosons of 1 : chemical potential µ, of the energy level εk at temperature T = kB β 1

nk =

eβ(εk −µ)

−1

(9.1)

As nk is positive or zero, the denominator exponent should be positive for any εk , which implies that the chemical potential µ must be smaller than any εk , including ε1 , the energy of the fundamental state. At ﬁxed T and µ, nk decreases when εk increases. The µ value is ﬁxed by the condition

nk = N

(9.2)

k

where N is the total particles number of the considered physical system. From nk one can calculate all the thermodynamical parameters and in particular the average energy : U=

εk nk

(9.3)

k

the grand potential A and its partial derivatives S and P A = −kB T ln(1 + nk )

(9.4)

k

dA = −P dΩ − SdT − N dµ

(9.5)

In the large volume limit, (9.1) is replaced by the Bose-Einstein distribution, a continuous function given by f (ε) =

1 eβ(ε−µ) − 1

(9.6)

The physical parameters are calculated in this case using the density of states D(ε), then expressing the constraint replacing (9.2) on the particles number and verifying that µ < ε1 . Here (§ 9.1) one is concerned by free bosons, with no p2 (apart from mutual interaction, therefore their energy is reduced to ε = 2m the “box” potential producing the quantization, leading to the expression of D(ε)), so that ε1 = 0 and µ < 0. Then for particles with a spin s, in the three-dimensional space, one gets,

Material Particles

203

Å ã √ 2s + 1 2m 3/2 taking D(ε) = CΩ ε, where C = [cf. Eq. (6.64)], 4π 2 2 √ ∞ ∞ εdε N= f (ε)D(ε)dε = CΩ β(ε−µ) e −1 0 0 ∞ ∞ 3/2 ε dε U= εf (ε)D(ε)dε = CΩ β(ε−µ) − 1 e 0 0 ∞ A = −kB T ln[1 + f (ε)]D(ε)dε 0 ∞ √ ε ln[1 − e−β(ε−µ) ]dε A = CΩkB T 2 = − CΩ 3

0 ∞

0

ε3/2 dε −1

eβ(ε−µ)

(9.7) (9.8)

(9.9) (9.10)

Integrating (9.9) by parts, one ﬁnds (9.10) which is equal to (9.8) to the factor 2 − , so that 3 2 A = − U = −P Ω (9.11) 3 This is the state equation of the bosons gas, that is, 2U (9.12) 3Ω which takes the same form as for the fermions (7.13), although the physical properties are totally diﬀerent. P =

9.1.2

Bose-Einstein Condensation

The above expressions are only valid when the condition µ < 0 is fulﬁlled (see above). Now when T decreases at constant density N/Ω, since β increases the diﬀerences (ε − µ) must decrease in order that N/Ω keeps its value (9.7). This means that µ gets closer to zero. Now consider the limit situation µ = 0. Relation (9.7) then becomes ∞ √ εdε N = CΩ βε − 1 e 0 √ Å ã CΩ ∞ xdx CΩ 1 N = 3/2 = 3/2 I x−1 e 2 0 βB βB

(9.13) (9.14)

i.e., using the mathematical tables of the present book, N=

CΩ 3/2

βB

· 2, 315

(9.15)

204

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Relation (9.15) deﬁnes a speciﬁc temperature TB = 1/kB βB which is only related to the density n = N/Ω : ã2/3 Å N 1 1 kB CΩ 2, 315 n2/3 6, 632 = 2mkB (2s + 1)2/3

TB =

In (9.17) C was replaced by its expression C =

2s + 1 4π 2

spin of the 42 He particles is zero.

(9.16) (9.17) Å

2m 2

ã3/2 , in which the

The temperature TB , or Bose temperature, would correspond in the above calculation to µ = 0, whereas the situation of § 9.1.1 is related to µ < 0. Since the integrals in (9.7) and (9.13) are both equal to N/CΩ, to maintain this value constant it is necessary that β (corresponding to µ < 0) be smaller than βB (associated to µ = 0). Consequently, (9.7) is only valid for T > TB . When lowering the temperature to TB , µ tends to zero, thus getting closer to the fundamental state of energy ε1 = 0. For T < TB the analysis has to be reconsidered : the population of level ε1 has to be studied separately, as it would diverge if one were using expression (9.1) without caution ! Now the population N1 of ε1 is at most equal to N . Expression (9.1) remains valid but µ cannot be strictly zero. Assume that N1 is a macroscopic number of the order of N : N1 =

1 1 = −βµ e −1 eβ(ε1 −µ) − 1

(9.18)

The corresponding value of µ is given by : 1 N1 kB T µ− N1

e−βµ = 1 +

(9.19)

µ is very close to zero because it is the ratio of a microscopic energy to a macroscopic number of particles. The ﬁrst excited level ε2 has the energy 2 ∝ Ω−2/3 , whereas the chemical potential is in Ω−1 , so that µ is very 2mL2 small as compared to ε2 . The N − N1 particles which are not located in ε1 are distributed among the excited states, for which one can take µ = 0 without introducing a large error,

Material Particles

205

and adapt the condition (9.7) on the particle number ∞ √ εdε N − N1 = CΩ β − 1 e ε 2∞ √ εdε CΩ βε − 1 e ε1 =0

(9.20) (9.21)

Indeed shifting the limit of the integral from ε2 to ε1 is adding a relative contribution to (9.20) of the order of (ε2 /kB T )1/2 . Since the volume Ω is macroscopic, this quantity is very small, ε2 being much less than kB T . One recognizes in (9.21) the integral (9.13) under the condition of replacing βB by β. Whence : N − N1 T 3/2 ñ Å ã ô T 3/2 N1 = N 1 − TB

N

3/2 TB

=

(9.22) (9.23)

Thus, when lowering the temperature, diﬀerent regimes take place : – as long as T > TB , the chemical potential is negative and the particles are distributed among all the microscopic states of the system according to the Bose-Einstein distribution (9.6) ; – when T = TB , µ almost vanishes (it is very small but negative) ; – when T < TB , the particles number N1 on the fundamental level becomes macroscopic, this is the “Bose condensation.” One thus realizes a macroscopic quantum state. The distribution of particles on the excited levels corresponds to µ 0. When T continues to decrease, N1 increases and tends to N when T tends to zero ; µ is practically zero. Going through TB corresponds to a phase transition : for T < TB the Bose gas is degenerate, which means that a macroscopic particles number is located in the same quantum state. This corresponds for 42 He to the superﬂuid state. The phenomenon of helium superﬂuidity is very complex and the interactions between atoms, neglected up to now, play a major part in this phase. The above calculation predicts a transition to the superﬂuid state at 3.2 K in standard conditions (4 He speciﬁc mass =140 kg/m3 under the atmospheric pressure) whereas the transition occurs at Tλ = 2.17 K. This is a second-order phase transition (the free energy is then ﬁnite, continuous and derivable ; the speciﬁc heat is discontinuous). The superconductivity of some solids also corresponds to a Bose condensation (2003 Nobel prize of A. Abrikosov and V. Ginzburg). It was possible to observe the atoms’ Bose condensation by trapping and cooling atomic beams at temperatures of the order of a mK in “optical molasses” (1997 Nobel prices of S. Chu, C. Cohen-Tannoudji and W.D. Phillips).

206

9.2

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Bose-Einstein Distribution of Photons

Our daily experience teaches us that a heated body radiates, as is the case for a radiator heating a room or an incandescent lamp illuminating us through its visible radiation. Understanding these mechanisms, a challenge at the end of the 19th century, was at the origin of M. Planck’s quanta hypothesis (1901). For more clarity we will not follow the historical approach but rather use the tools provided by this course. Once the results obtained, we will return to the problems Planck was facing.

9.2.1

Description of the Thermal Radiation ; the Photons

Consider a cavity of volume Ω heated at temperature T (for example an oven) and previously evacuated. The material walls of this cavity are made of atoms, the electrons of which are promoted into excited states by the energy from the heat source. During their deexcitation, these electrons radiate energy, under the form of an electromagnetic ﬁeld which propagates inside the cavity, is absorbed by other electrons associated to other atoms, and so forth, and this process leads to a thermal equilibrium between the walls and the radiation enclosed inside the cavity. If the cavity walls are perfect reﬂectors, classical electromagnetism teaches us is zero inside the perfect conductor and normal that the wave electrical ﬁeld E tangential is contito the wall on the vacuum side (due to the condition that E nuous), and that the corresponding magnetic ﬁeld B must be tangential. In order to produce the resonance of this cavity and to establish stationary waves, the cavity dimension L should contain an integer number of half-wavelengths λ. The wave vector k = 2π/λ and L are then linked. k) exp i(k ·r −ωt) An electromagnetic wave, given in complex notation by E(ω, is characterized by its wave vector k and its frequency ω ; it can have two independent polarization states (either linear polarizations along two perpendicular directions, or left or right circular polarizations) on which the direction of the electric ﬁeld in the plane normal to k is projected. Finally its intensity k)|2 . is proportional to |E(ω, Since the interpretation of the photoelectric eﬀect by A. Einstein in 1905, it has been known that the light can be described both in terms of waves and of particles, which are the photons. The quantum description of the electromagnetic ﬁeld is taught in advanced Quantum Mechanics courses. Here we only specify that the photon is a relativistic particle of zero mass, so that

Bose-Einstein Distribution of Photons

207

the general relativistic relation between the energy ε and the momentum p, ε = p2 c2 + m20 c4 , where c is the light velocity in vacuum, reduces in this case to ε = pc. Although the photon spin is equal to 1, due to its zero mass it only has two distinct spin states. The set of data consisting in p and the photon spin value deﬁnes a mode. The wave parameters k and ω and the photon parameters are related through the Planck constant, as summarized in this table : electromagnetic wave wave vector k frequency ω k, ω)|2 intensity |E( polarization (2 states)

particle : photon p = k momentum ε = ω energy number of photons spin (2 values)

Let us return to the cavity in thermal equilibrium with the radiation it contains : its dimensions, in Ω1/3 , are very large with respect to the wavelength of the considered radiation. The exact surface conditions do not matter much : as in the particles case (see § 6.4.1b) ), rather than using the stationary wave conditions, we prefer the periodic limit conditions (“Born-Von Kármán” conditions) : in a thought experiment we close the system on itself, which quantizes k and consequently, p and the energy. Assuming the cavity is a box of dimensions Lx , Ly , Lz , one obtains : ã Å k = (kx , ky , kz ) = 2π nx , ny , nz (9.24) Lx Ly Lz where nx , ny , nz are positive or negative integers, or zero. Then

Å p = h

nx ny nz , , Lx Ly Lz

ε = pc = ω = hν

9.2.2

ã (9.25) (9.26)

Statistics of Photons, Bosons in Non-Conserved Number

In the photon emission and absorption processes of the wall atoms, the number of photons varies, and there are more photons when the walls are hotter : in this system the constraint on the conservation of particles number, that existed in § 9.1, is lifted. This constraint was deﬁning the chemical potential µ as a Lagrange multiplier, which no longer appears in the probability law for photons, i.e., the chemical potential is zero for photons.

208

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Let us consider the special case where only a single photon mode is possible, i.e., a single value of p, of the polarization and of ε in the cavity at thermal equilibrium at temperature T . Any number of photons is possible, thus the partition function z(ε) is equal to z(ε) =

∞

e−βnε =

n=0

1 1 − e−βε

(9.27)

and the average energy at T of the photons in this mode is given by nε = −

ε 1 ∂z(ε) = βε z(ε) ∂β e −1

For the occupation factor of this mode, one ﬁnds again the Bose-Einstein distribution in which µ = 0, i.e. f (ε) =

1 −1

(9.28)

eβε

[Note that in the study of the speciﬁc heat of solids (see § 2.4.5) a factor similar to (9.28) appeared : in the Einstein and Debye models the quantized solid vibrations are described by oscillators. Changing the vibration

harmonic

state of the solid from n + 12 ω to n + 12 + 1 ω is equivalent to creating a quasi-particle, called a phonon, which follows the Bose-Einstein statistics. The number of phonons is not conserved]. If now one accounts for all the photons modes, for a macroscopic cavity volume p), D(ε) such that the Ω, one can deﬁne the densities of states Dk (k), Dp ( number dn of modes with a wave vector between k and k + dk is equal to dn : Ω 3 d k = Dk (k)d3k (2π)3 Ω p = Dp ( p)d3 p dn = 2 · 3 d3 h dn = 2 ·

that is, Dk (k) =

Ω 4π 3

,

Dp ( p) =

2Ω h3

(9.29)

The factor 2 expresses the two possible values of the spin of a photon of ﬁxed k (or p). Since the energy ε only depends of the modulus p, one obtains D(ε) from Ω Ω 4πp2 dp = 8π 3 3 ε2 dε 3 h h c 8πΩ 2 D(ε) = 3 3 ε h c

D(ε)dε = 2

(9.30)

Bose-Einstein Distribution of Photons

209

In this “large volume limit”, one can express the thermodynamical parameters. The thermodynamical potential to be considered here is the free energy F (since µ = 0) ∞ ln(1 − e−βε )D(ε)dε (9.31) F = −kB T ln Z = kB T 0

The internal energy is written ∞ U= εD(ε)f (ε)dε 0 ∞ ∞ 8πΩ ε3 8πΩ hν 3 dν = dε = h3 c3 eβε − 1 c3 eβhν − 1 0 0 i.e.,

(9.32)

∞

Ωu(ν)dν

U= 0

with u(ν) =

8πh ν3 c3 eβhν − 1

(9.33)

The photon spectral density in energy u(ν) is deﬁned from the energy contribution dU of the photons present inside the volume Ω, with a frequency between ν and ν + dν : dU = Ωu(ν)dν

(9.34)

The expression (9.33) of u(ν) is called the Planck law . The total energy for the whole spectrum is given by ∞ 3 8πΩ x dx U = 3 3 (kB T )4 h c ex − 1 0

(9.35)

where the dimensionless integral is equal to Γ(4)ζ(4) with Γ(4) = 3! and π4 ζ(4) = (see the section “Some useful formulae” in this book). Introducing 90 = h/2π, the total energy becomes U=

π 2 (kB T )4 Ω 15 (c)3

(9.36)

The free energy U and the internal energy F are easily related : in the integration by parts of (9.31), the derivation of the logarithm introduces

210

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

the Bose-Einstein distribution, the integration of D(ε) provides a term in ε3 /3 = ε . ε2 /3. This exactly gives F =−

π 2 (kB T )4 U =− Ω 3 45 (c)3

[to be compared to (9.11), which is valid for material particles with a density of states in ε1/2 , whereas the photons density varies in ε2 ]. Since dF = −SdT − P dΩ

(9.37)

one deduces Å S=− Å P =− P =

∂F ∂T ∂F ∂Ω

ã = Ω

ã

Å ã 4π 2 kB T 3 kB Ω 45 c

=− T

F Ω

π 2 (kB T )4 45 (c)3

(9.38) (9.39)

(9.40)

The pressure P created by the photons is called the radiation pressure.

9.2.3

Black Body Deﬁnition and Spectrum

The results obtained above are valid for a closed cavity, in which a measurement is strictly impossible because there is no access for a sensor. One is easily convinced that drilling a small hole into the cavity will not perturb the photons distribution but will permit measurements of the enclosed thermal radiation, through the observation of the radiation emitted by the hole. Besides any radiation coming from outside and passing through this hole will be trapped inside the cavity and will only get out after thermalization. This system is thus a perfect absorber, whence its name of black body. In addition we considered that thermal equilibrium at temperature T is reached between the photons and the cavity. In such a system, at steady-state, for each frequency interval dν, the energy absorbed by the walls exactly balances their emitted energy, which is found in the photon gas in this frequency range, that is, Ωu(ν)dν. Therefore the parameters we obtained previously for the photon gas are also those characterizing the thermal emission from matter at temperature T , whether this emission takes place in a cavity or not, whether this matter is in contact with vacuum or not. We will then be able to compare the laws already stated, and in the ﬁrst place the Planck law, to experiments performed on thermal radiation.

Bose-Einstein Distribution of Photons

211

The expression (9.33) of u(ν) provides the spectral distribution of the thermal radiation versus the parameter β, i.e. versus the temperature T . For any T , u(ν) vanishes for ν = 0 and tends to zero when ν tends to inﬁnity. The variable is in fact βhν and one ﬁnds that the maximum of u(ν) is reached for βhνmax = 2.82, that is, νmax = 2.82

kB T h

(9.41)

Relation (9.41) constitutes the Wien law. When T increases, the frequency of the maximum increases and for T1 < T2 the whole curve u(ν) for T2 is above that for T1 (Fig. 9.1). u(ν)

J.s.m−3

5.10−15 T1 < T2 < T3

T2

T3

=

60

00

K

T1 0

ν 10 Hz 14

4

Fig. 9.1: Spectral density in energy of the thermal radiation for several temperatures.

For small values of ν, u(ν) kB T · for ν large, u(ν)

8π 2 ν c3

8πh 3 −βhν ν e c3

(9.42) (9.43)

At the end of the 19th century, photometric measures allowed one to obtain the u(ν) curves with a high accuracy. Using the then available MaxwellBoltzmann statistics, Lord Rayleigh and James Jeans had predicted that the average at temperature T of the energy of oscillators should take the form (9.42), in kB T ν 2 . If this latter law does describe the low-frequency behavior of the thermal radiation, it predicted an “ultraviolet catastrophe” [u(ν) would be an increasing function of ν, thus the U V and X emissions should be huge]. Besides, theoretical laws had been obtained by Wilhem Wien and experimentally conﬁrmed by Friedrich Paschen : an exponential behavior at

212

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

high frequency like in (9.43) and the law (9.41) of the displacement of the distribution maximum. It was the contribution of Max Planck in 1900 to guess the expression (9.33) of u(ν), assuming that the exchanges of energy between matter and radiation can only occur through discrete quantities, the quanta. Note that the ideas only slowly clariﬁed until the advent in the mid 1920’s of Quantum Mechanics, such as we know it now. It it remarkable that the paper by Bose, proposing the now called “Bose-Einstein statistics,” preceded by a few months the formulation by Erwin Schroedinger of its equation ! The universe is immersed within an infrared radiation, studied in cosmology, the distribution of which very accurately follows a Planck law for T = 2.7 K. It is in fact a “fossil” radiation resulting from the cooling, by adiabatic expansion of the universe, of a thermal radiation at a temperature of 3000 K which was prevailing billions of years ago : according to expression (9.38) of the entropy, an adiabatic process maintains the product ΩT 3 . The universe radius has thus increased of a factor 1000 during this period. The sun is emitting toward us radiation with a maximum in the very near infrared, corresponding to T close to 6000 K. A “halogen” bulb lamp is emitting a radiation corresponding to the temperature of its tungsten ﬁlament (around 3000 K), which is immersed in a halogen gas to prevent its evaporation. We ourselves are emitting radiation with a maximum in the far infrared, corresponding to our temperature close to 300 K.

9.2.4

Microscopic Interpretation of the Bose-Einstein Distribution of Photons

Another schematic way to obtain the Bose-Einstein distribution, and consequently, the Planck law, consists in considering that the walls of the container are made of atoms with only two possible energy levels ε0 and ε1 separated by ε1 − ε0 = hν, and in assuming that the combined system (atoms-photons of energy hν) is at thermal equilibrium (Fig. 9.2). It is possible to rigorously calculate the transition probabilities between ε0 and ε1 (absorption), or between ε1 and ε0 (emission) under the eﬀect of photons of energy hν. This is done in advanced Quantum Mechanics courses. Here we will limit ourselves to a more qualitative argument, admitting that in Quantum Mechanics a transition is expressed by a matrix element, which is the product of a term 1|V |0 speciﬁc of the atoms with another term due to the photons. Besides, we will admit that an assembly of n photons of energy hν is described by a harmonic oscillator, with energy levels splitting equal to hν, lying in the state |n. The initial state contains N0 atoms in the fundamental state, N1 atoms in the state ε1 , and n photons of energy hν.

Bose-Einstein Distribution of Photons

213

BEFORE

AFTER

N1 hν

E1

N0

E0

N1

E1

N0

E0

N1

E1

N1 + 1 absorption N0 − 1

N1 − 1 hν spontaneous emission N0 + 1

N1 − 1 hν

hν N0

E0

induced emission

N0 + 1

Fig. 9.2: Absorption, spontaneous and induced emission processes for an assembly of two-level atoms in thermal equilibrium with photons.

In the absorption process (Fig. 9.2 top), the number of atoms in the state of energy ε0 becomes N0 − 1, while N1 changes to N1 + 1 ; a photon is utilized, thus the photons assembly shifts from |n to |n − 1. This is equivalent to considering the action on the state |n of the photon annihilation operator, i.e., according to the Quantum Mechanics course, a|n =

√ n |n − 1

(9.44)

The absorption probability per unit time P is proportional to the number of atoms N0 capable of absorbing a photon, to the square of the matrix element appearing in the transition, and is equal, to a multiplicative constant, to : Pa = N0 |1|V |0|2 |n − 1|a|n|2 = N0 |1|V |0| n 2

(9.45) (9.46)

This probability is, as expected, proportional to the number n of present photons.

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Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

In the emission process (Figs. 9.2 middle and bottom), using a similar argument, since the number of photons is changed from n to n + 1, one considers the action on the state |n of the photon creation operator : √ (9.47) a+ |n = n + 1 |n + 1 The emission probability per unit time Pe is proportional to the number of atoms N1 that can emit a photon, to the square of the matrix element of this transition and is given by Pe = N1 |0|V |1|2 |n + 1|a+ |n|2 = N1 |0|V |1| (n + 1) 2

(9.48) (9.49)

We ﬁnd here that Pe is not zero even when n = 0 ( spontaneous emission), but that the presence of photons increases Pe (induced or stimulated emission) : another way to express this property is to say that the presence of photons “stimulates” the emission of other photons. The induced emission was introduced by A. Einstein in 1916 and is at the origin of the laser eﬀect. Remember that the name “laser” stands for Light Ampliﬁcation by Stimulated Emission of Radiation. In equilibrium at the temperature T , n photons of energy hν are present : the absorption and emission processes exactly compensate, thus N0 n = N1 (n + 1)

(9.50)

Besides, for the atoms in equilibrium at T , the Maxwell-Boltzmann statistics allows one to write N1 = e−βhν (9.51) N0 One then deduces the value of n n=

1 eβhν

−1

(9.52)

which is indeed the Bose-Einstein distribution of photons. This yields u(ν, T ) through the same calculation of the photons modes as above [formulae (9.29) to (9.33)]. This approach, which is based on advanced Quantum Mechanical results, has the advantage of introducing the stimulated emission concept and the description of an equilibrium from a balance between two phenomena.

9.2.5

Photometric Measurements : Deﬁnitions

We saw that a “black body” is emitting radiation corresponding to the thermal equilibrium in the cavity at the temperature T . A few simple geometrical

Bose-Einstein Distribution of Photons

215

considerations allow us to relate the total power emitted by a black body to its temperature and to introduce the luminance, which expresses the “color” of a heated body : we will then be able to justify that all black bodies at the same given temperature T are characterized by the same physical parameter and thus that one can speak of THE black body. Let us ﬁrst calculate the total power P emitted through the black-body aperture in the half-space outside the cavity. We consider the photons of frequency ν, to dν, which travel through the aperture of surface dS and have a velocity directed toward the angle θ to dω with the normal n to dS. The number of these photons which pass the hole during dt are included in a cylinder of basis dS and height c cos θdt (Fig. 9.3). They carry an energy

dω θ dS

n

cdt

Fig. 9.3: Radiation emitted through an aperture of surface dS in a cone dω around the direction θ.

dω c cos θ dt dS u(ν)dν (9.53) 4π The total power P radiated into the half-space, by unit time, through a hole of unit surface is given by 2π π2 dϕ ∞ cos θ sin θdθ c u(ν, T )dν (9.54) P = 4π 0 0 0 d2 P · dt =

P =

π 2 (kB T )4 cU = = σT 4 4Ω 60 3 c2

(9.55)

This power varies like T 4 , this is the Stefan-Boltzmann law. The Stefan constant σ is equal to 5.67 × 10−8 watts · m−2 · kelvins−4 . At 6000 K, the approximate temperature of the sun surface, one gets P = 7.3 × 107 W· m−2 ; at 300 K one ﬁnds 456 W·m−2 . Another way to handle the question is to analyze the collection process of the radiation issued from an arbitrary source by a detector. It is easy to

216

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

understand that the collected signal ﬁrst depends on the spectral response of the detector : for instance, a very speciﬁc detector is our eye, which is only sensitive between 0.4 µm and 0.70 µm, with a maximum of sensitivity in the yellow-green at 0.55 µm. Therefore speciﬁc units have been deﬁned in visual photometry, which take into account the physiological spectral response. But the signal also depends on geometrical parameters, such as the detector surface dS , the solid angle under which the detector sees the source or, equivalently, the source surface dS and the source-detector distance together with their respective directions (Fig. 9.4). A physical parameter, the luminance L, characterizes the source : in the visible detection range the colors of two sources of diﬀerent luminances are seen as diﬀerent (a red-hot iron does not have the same aspect as the sun or as an incandescent lamp ; a halogen lamp looks whiter that an electric torch). To be more quantitative, in energetic photometry one deﬁnes the inﬁnitesimal power d2 P received by the elementary surface dS of a detector from the surface dS of a source (Fig. 9.4) by the equation d2 P = LdS cos θdωdν = LdS

cos θ cos θ dS dν = LdS cos θ dω dν r2

dS

dS dω M

(9.56)

θ n

dω θ

M

n

Fig. 9.4: Deﬁnition of the geometrical parameters of the source and detector. The surface element of the source is centered on M , the detection element on M , with M M = r ; θ is the angle between M M and the normal n to dS, dω the solid angle under which the detector is seen from the source dS. One notices that the deﬁnition is fully symmetrical and that alternatively one can introduce dS , the solid angle dω under which the source is seen from the detector, the angle θ between M M and the normal to dS in M . The luminance depends of T , ν and possibly of θ, the direction of the emission, that is, L(ν, T, θ). When the source is a black body of surface dS in thermal equilibrium at the temperature T , the power emitted toward dS in the frequency range (ν, ν+dν)

Bose-Einstein Distribution of Photons

217

is calculated from (9.53) d2 P =

c u(ν, T ) dS cos θ dω dν 4π

(9.57)

The comparison between (9.56) and (9.57) allows one to express the black body luminance L0 versus u(ν, T ) : c u(ν, T ) (9.58) L0 (ν, T ) = 4π Thus all black bodies at the same temperature T have the same luminance, which allows one to deﬁne the black body. It is both the ideal source and detector since, once the equilibrium has been reached, it absorbs any radiation that it receives and re-emits it. Its luminance does not depend on the direction θ of emission. Formulae (9.56) and (9.57) also express that the surface which comes into play is the apparent surface, i.e., the projection dS cos θ of the emitting surface on the plane perpendicular to M M . For this whole surface c u(ν, T ) : this is the Lambert law the luminance is uniform and equal to 4π which explains why the sun, a sphere radiating like the black body, appears to us like a planar disk. The heated bodies around us are generally not strictly black bodies, in the meaning that (9.58) is not fulﬁlled : their luminance(L(ν, T, θ)) not only depends on the frequency and the temperature, but also on the emission direction. Besides, we saw that the black body is “trapping” any radiation which enters the cavity : it thus behaves like a perfect absorber, which is not the most general case for an arbitrary body.

9.2.6

Radiative Balances

Consider an arbitrary body, exchanging energy only through radiation and having reached thermal equilibrium : one can only express a balance stating that the absorbed power is equal to the emitted power (in particular a transparent body neither absorbs nor emits energy). Assume that the studied body is illuminated by a black body and has an absorption coeﬃcient a(ν, T, θ) (equal to unity for the black body). If the studied body does not diﬀuse light, that is, radiation is only reemitted in the image direction of the source, and if it remains at the frequency it was absorbed, one obtains the relation c u(ν, T )a(ν, T, θ) = L(ν, T, −θ) (9.59) L0 (ν, T )a(ν, T, θ) = 4π emission absorption after simplifying the geometrical factors. Relation (9.59) links the black body luminance L0 (ν, T ) to that of the considered body, of absorption coeﬃcient a(ν, T, θ). The black body luminance is

218

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

larger than, or equal to, that of any thermal emitter at the same temperature, the black body is the best thermal emitter. This is the Kirchhoﬀ law. In the case of a diﬀusing body (reemitting radiation in a direction other than −θ) or of a ﬂuorescent body (reemitting at a frequency diﬀerent from that of its excitation), one can no longer express a detailed balance, angle by angle or frequency by frequency, between the ambient radiation and the body radiation. However, in steady-state regime, the total absorbed power is equal to the emitted power and the body temperature remains constant. The applicable relation is deduced from (9.59) after integration over the angles and/or over the frequencies. One can also say that the power, assumed to be emitted by a black body, received by the studied body at thermal equilibrium at T , is either absorbed (and thus reemitted since T remains constant), or reﬂected. For instance for a diﬀusing body one writes : received power = emitted power + reﬂected power which implies for the luminances : L0 (ν, T ) = L(ν, T ) + [L0 (ν, T ) − L(ν, T )] L0 (ν, T ) = L(ν, T ) + [1 − a(ν, T )]L0 (ν, T )

(9.60)

This is another way to obtain (9.59). These types of balances are comparisons between the body luminance L and the black body luminance L0 : if the interpretation of the L0 expression requires Quantum Mechanics, the comparisons of luminances are only based on balances and were discovered during the 19th century. More generally speaking, the exchanges of energy are taking place according to three processes : conduction, convection and radiation. Whereas the ﬁrst two mechanisms require a material support, radiation which occurs through photons can propagate through vacuum : it is thanks to radiation that we are in contact with the universe. Through the intersideral vacuum, stars are illuminating us and we can study them by analyzing the received radiation. Our star, the sun, is sending a power of the order of a kW to each 1 m2 surface of earth. The researches on solar energy try to make the best use of this radiation, using either thermal sensors, or photovoltaic cells (“solar cells”) (see §8.5.2) which directly convert it into electricity.

9.2.7

Greenhouse Eﬀect

Before interpreting the radiative balance of the earth, let us consider the greenhouse eﬀect which allows one to maintain inside a building, made of

Bose-Einstein Distribution of Photons

219

glass formerly, often of plastics nowadays, a temperature higher than the one outside and thus favors particular vegetables or ﬂowers cultivation (Fig. 9.5). The origin of the eﬀect arises from the diﬀerence in absorption properties of the building walls for the visible and infrared radiations. In the visible range, where the largest part of the radiation emitted by the sun is found, the walls are transparent, so that the sun radiation can reach the ground, where it is absorbed. The ground in turn radiates like a black body, with its frequency maximum in the infrared [Wien law, equation (9.41)]. Now the walls are absorbing in the infrared, they reemit a part of this radiation toward the ground, the temperature of which rises, and so on. At equilibrium the ground temperature is thus TG. If a steady state has been reached, the power getting into the greenhouse, i.e., P0 , is equal to the one that is getting out of it :

P0

Wall in glass or plastics a

TG

Ground TG

Fig. 9.5: Principle of the greenhouse eﬀect.

P0 = σSTG4 (1 − a) if S is the ground surface and a the absorption coeﬃcient, by the walls, of the ground infrared radiation. In the absence of greenhouse, the ground temperature would be TG , such that P0 = σSTG4 The temperature TG inside the greenhouse is higher than TG since TG =

TG (1 − a)1/4

Here we assumed a perfect transparency of the walls in the visible and no reﬂection by the walls both in the visible and the infrared. “The greenhouse eﬀect” related to a possible climate global warming is an example of radiative balance applied to the earth. The earth is surrounded by a layer of gases, some of which corresponding to molecules with a permanent dipolar moment, due to their unsymetrical arrangement of atoms : this is the case for H2 O, CO2 , O3 , etc. Thus the vibration and rotation spectra of these molecules evidence an infrared absorption. The sun illuminates the earth with thermal radiation, the maximum of its spectrum being located in the visible [Eq. (9.41)] ; this radiation is warming

220

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

the ground. The earth, of temperature around 300 K, reemits thermal radiation toward space, the frequency maximum of this reemission is in the infrared. A fraction of the earth radiation is absorbed by the gases around the earth, which in turn reemit a fraction of it, and so on. Finally, according to the same phenomenon as the one described above for a greenhouse, the earth receives a power per unit surface larger than the one it would have reached it if there had only been vacuum around it ; its average temperature is 288 K (instead of 255 K if it had been surrounded by vacuum). The global climate change, possibly induced by human activity, is related to an increase in concentration of some greenhouse eﬀects gases (absorbing ﬂuorocarboned compounds, which are now forbidden for use, and mainly CO2 ). If the absorption coeﬃcient increases, the ground temperature will vary in the same way. This will produce eﬀects on the oceans level, etc. . . 1 More advanced studies on the greenhouse eﬀect account for the spectra of absorption or transparency of the gases layer around the earth according to their chemical species, etc. In our daily life, we are surrounded by radiating objects like heating radiators, incandescent lamps. On the contrary we wish other objects to maintain their energy without radiating : this is the case, in particular, for thermos bottles, Dewar cans. We foresee here a very vast domain with technological, ecological and political involvements !

1 The Intergovernmental Panel on Climate Change, established by the World Meteorological Organization and the United Nations Environment program, is assessing the problem of potential global climate change. An introduction brochure can be downloaded from its website (www.ipcc.ch).

Summary of Chapter 9 The chemical potential µ of material non-interacting particles following the Bose-Einstein statistics is located below their fundamental state ε1 . The occupation factor by material bosons of an energy level ε is equal to f (ε) =

1 eβ(ε−µ) − 1

At given density N/Ω, below a temperature TB , which is a function of N/Ω, the Bose condensation takes place : a macroscopic number of particles then occupies the fundamental state ε1 , that is, the same quantum state. The thermodynamical properties in this low temperature range are mainly due to the particles still in the excited states. The superﬂuidity of helium 4, the superconducting state of some solids, are examples of systems in the condensed Bose state, which is observed for T < TB . Photons are bosons in non-conserved number, their Bose-Einstein distribution is given by 1 f (ε) = βε e −1 The Planck law expresses the spectral density in energy or frequency of photons in an empty cavity of volume Ω, in the frequencies range [ν, ν + dν] : dU = Ωu(ν)dν = Ω

8πh ν 3 dν c3 eβhν − 1

This law gives the spectral distribution of the “black body” radiation, i.e., of thermal radiation at temperature T = 1/kB β. The black body is the perfect thermal emitter, its luminance is proportional to u(ν). The thermal emission of a body heated at temperature T corresponds to a luminance at maximum equal to that of the black body. 221

222

Summary of Chapter 9

The greenhouse eﬀect arises from the properties of the walls enclosing the greenhouse : higher transparency in the visible and larger infrared absorption. This phenomenon is also applicable to the radiation balance of the earth.

General Method for Solving Exercises and Problems At the risk of stating obvious facts, we recall here that the text of a statistical physics exercise or problem is very precise, contains all the required information and that each word really matters. Therefore it is recommended to ﬁrst read the text with the highest attention. The logic of the present course will always be apparent, i.e, you have to ask yourself the two following questions, in the order given here : 1. What are the microstates of the studied system ? 2. Which among these microstates are achieved in the problem conditions ? 1. The ﬁrst question is generally solved through Quantum Mechanics, which reduces to Classical Mechanics in special cases (the ideal gas, for example). Mostly this is equivalent to the question : “What are the energy states of the system ?” In fact they are often speciﬁed in the text. At this stage there are three cases : – distinguishable particles (for example of ﬁxed coordinates) : see the ﬁrst part of the course – classical mobile particles : see the chapter on the ideal gas – indistinguishable particles : see the second part of the course, in which the study is limited to the case of independent particles.

223

224

General Method for Solving Exercises and Problems

2. The second question speciﬁcally refers to Statistical Physics. According to the physical conditions of the problem one will work in the adapted statistical ensemble : – isolated system, ﬁxed energy and ﬁxed particles number : microcanonical ensemble – temperature given by an energy reservoir and ﬁxed number of particles : canonical ensemble – system in contact with an energy reservoir (which dictates its temperature) and a particles reservoir (which dictates its chemical potential) : grand canonical ensemble The systems of indistinguishable particles obeying the Pauli principle and studied in the framework of this course are solved in the grand-canonical ensemble. Their chemical potential is then set in such a way that the average number of particles coincides with the real number of particles given by the physics of the system.

Units and physical constants (From R. Balian, “From microphysics to macrophysics : methods and applications of statistical physics”, Springer Verlag Berlin (1991)) We use the international system (SI), legal in most countries and adopted by most international institutions. Its fundamental units are the meter (m), the kilogram (kg), the second (sec), the ampere (A), the kelvin (K), the mole (mol), and the candela (cd), to which we add the radian (rad) and the steradian (sr). The derived units called by speciﬁc names are : the hertz (Hz = s−1 ), the newton (N = m kg sec−2 ), the pascal (Pa = N m−2 ), the joule (J = N m), the watt (W = J sec−1 ), the coulomb (C = A s), the volt (V = W A−1 ), the farad (F = C V−1 ), the ohm (Ω = V A−1 ), the siemens (S = A V−1 ), the weber (Wb = V s), the tesla (T = Wb m−2 ), the henry (H = Wb A−1 ), the Celsius degree (C), the lumen (lm = cd sr), the lux (lx = lm m−2 ), the becquerel (Bq = sec−1 ), the gray (Gy = J kg−1 ) and the sievert (Sv = J kg−1 ). The multiples and submultiples are indicated by the preﬁxes deca (da = 10), hecto (h = 102 ), kilo (k = 103 ) ; mega (M = 106 ), giga (G = 109 ), tera (T = 1012), peta (P = 1015 ), exa (E = 1018 ) ; deci (d = 10−1 ), centi (c = 10−2 ), milli (m = 10−3 ) ; micro (µ = 10−6 ), nano (n = 10−9 ), pico (p = 10−12 ), femto (f = 10−15 ), atto (a = 10−18 ).

225

226

Units and physical constants

Constants of electromagnetic units light velocity Planck constant Dirac constant Avogadro number Atomic mass unit neutron and proton masses electron mass Elementary charge Faraday constant Bohr magneton nuclear magneton Fine structure constant Hydrogen atom : Bohr radius binding energy Rydberg constant Boltzmann constant molar constant of the gas Normal conditions : pressure temperature

molar volume Gravitation constant gravity acceleration Stefan constant

µ0 = 4π × 10−7 N A−2 (deﬁnes the ampere) ε0 = 1/µ0 c2 , 1/4πε0 9 × 109 N m2 C−2 c = 299 792 458 m sec−1 (deﬁnes the meter) c 3 × 108 m sec−1 h = 6.626 × 10−34 J sec = h/2π 1.055 × 10−34 J sec N 6.022 × 1023 mol−1 (by deﬁnition, the mass of one mole of 12 C is 12 g) 1 u=1 g/N 1.66 × 10−27 kg (or dalton or amu) mn 1.0014mp 1.0088 u m 1 u/1823 9.11 × 10−31 kg e 1.602 × 10−19 C N e 96 485 C mol−1 µB = e/2m 9.27 × 10−24 J T−1 e/2mp 5 × 10−27 J T−1 e2 1 α= 4πε0 c 137 4πε0 2 = 0.53 Å mcα me2Å ã2 2 m e2 E0 = = 2 2ma20 2 4πε0 13.6 eV R∞ = E0 /hc 109 737 cm−1 kB 1.381 × 10−23 J K−1 R = N kB 8.316 J K−1 mol−1 a0 =

1 atm = 760 torr = 1.01325× 105 Pa Triple point of water : 273.16 K ( deﬁnition of the kelvin) or 0.01C (deﬁnition of the Celsius scale) 22.4 × 10−3 m3 mol−1 G 6.67 × 10−11 m3 kg−1 sec−2 g 9.81 m sec−2 π2 k4 603 c2 5.67 × 10−8 W m−2 K−4 σ=

Units and physical constants

227

Deﬁnition of the photometric units

A light power of 1 W, at the frequency of 540 THz, is equivalent to 683 lm

Energy units and equivalences :

1 erg = 10−7 J (not IS) 1 kWh=3.6×106 J 1 eV ↔ 1.602 × 10−19 J ↔ 11 600 K

electrical potential (electron-volt) heat (calorie) chemical binding temperature (kB T ) mass (mc2 ) wave number (hc/λ) frequency (hν)

1 cal = 4.184 J (not SI ; speciﬁc heat of 1 g water) 23 kcal mol−1 ↔ 1 eV (not SI) 1 290 K ↔ 40 eV (standard temperature) 9.11 × 10−31 kg ↔ 0.511 MeV (electron at rest) 109 700 cm−1 ↔ 13.6 eV (Rydberg) 3.3 × 1015 Hz ↔ 13.6 eV

It is useful to remember these equivalencies to quickly estimate orders of magnitude. Various non-SI units

1 ångström (Å)=10−10 m (atomic scale) 1 fermi (fm)=10−15 m (nuclear scale) 1 barn (b)=10−28 m2 1 bar=105 Pa 1 gauss (G)=10−4 T 1 marine mile=1852 m 1 knot=1 marine mile per hour 0.51 m sec−1 1 astronomical unit (AU) 1.5 × 1011 m (Earth-Sun distance) 1 parsec (pc) 3.1 × 1016 m (I UA/sec arc) 1 light-year 0.95 × 1016 m

Data on the Sun

Radius 7 × 108 m=109 earth radii Mass 2 × 1030 kg Average density 1.4 g cm−3 Luminosity 3.8 × 1026 W

A few useful formulae Normalization of a Gaussian function : … +∞ π −ax2 dx e = a −∞ The derivation of this formula with respect to a yields the moments of the Gauss distribution.

Euler gamma function :

∞

Γ(t) ≡

xt−1 e−x dx = (t − 1)Γ(t − 1)

0

∞

(1 − x)s−1 xt−1 dx =

0

Γ(t)Γ(1 − t) = Stirling formula :

π sin πt

,

Γ(s)Γ(t) Γ(s + t)

Γ

Å ã √ 1 = π 2

√ t! = Γ(t + 1) ∼ tt e−t 2πt t→∞

Binomial series : (1 + x)t =

∞ n x

∞ (−x)n Γ(n − t) Γ(t + 1) = n! Γ(t + 1) − n) n=0 n! Γ(−t) n=0

,

Poisson formula : +∞

f (n) =

n=−∞

+∞

f˜(2πl) ≡

l=−∞

+∞ l=−∞

229

+∞

−∞

dxf (x)e2πilx

|x| < 1

230

A few useful formulae

Euler-Maclaurin formula : ε 1 ε3 1 a+ε f (x) |a+ε dxf (x) ≈ [f (a) + f (a + ε)] − f (x) |a+ε + +... a a ε a 2 12 720 ε 7ε3 ≈ f (a + 12 ε) + f (x) |a+ε f (x) |a+ε − +... a a 24 5760 This formula allows one to calculate the diﬀerence between an integral and a sum over n, for a = nε. Constants e 2.718, π 3.1416 1 γ ≡ lim(1 + ... + − ln n) 0.577 Euler constant. n Riemann zeta function : ∞ t−1 ∞ 1 x dx = Γ(t)ζ(t) ζ(t) ≡ , t n ex − 1 0 n=1

∞

0

t ζ

1.5 2.612

2

xt−1 dx = (1 − 2−t+1 )Γ(t)ζ(t) ex + 1

2.5 1.341

1 2 6π

Dirac distribution : 1 2π

+∞

−∞

3 1.202

3.5 1.127

4 1 4 90 π

5 1.037

dx eixy/a = δ(y/a) = |a|δ(y)

+∞ 1 2πi y/a e = δ(y/a) |a|

=−∞

lim

t→∞

sin tx 1 − cos tx = lim = πδ(x) t→∞ x tx2

f (x)δ(x) = f (0)δ(x)

,

f (x)δ (x) = −f (0)δ(x) + f (0)δ (x)

If f (x) = 0 at the x = xi points, one has 1 δ(x − xi ) δ[f (x)] = |f (xi )| i

Exercises and Problems A course can only be fully understood after practice through exercises and problems. There are many excellent collections of classical exercises in Statistical Physics. In what follows, we reproduce original texts which were recently given as examinations at Ecole Polytechnique. An examination of Statistical Physics at Ecole Polytechnique usually consists in an exercise and a problem. As examples, you will ﬁnd here the exercises given in 2000, 2001 and 2002 and the problems of 2001 and 2002 which were to be solved by the students who had attended the course presented in this book. Electrostatic Screening : exercise 2000 Magnetic Susceptibility of a “Quasi-1D” Conductor : exercise 2001 Entropies of the HC Molecule : exercise 2002 Quantum Boxes and Optoelectronics : problem 2001 Physical Foundations of Spintronics : problem 2002

231

Exercise 2000 : Electrostatic Screening

233

Exercise 2000 : Electrostatic Screening I. Let us consider N free fermions of mass me and spin 1/2, without any mutual interaction, contained in a macroscopic volume Ω maintained at temperature T . I.1. Recall the equation relating N/Ω to the chemical potential µ of these fermions through the density of states in k (do not calculate the integral). I.2. In the high temperature and low density limit recall the approximate integral expression relating N/Ω to µ (now calculate the integral). I.3. The same fermions, at high temperature and small density, are now submitted to a potential energy V (r). Express the total energy of one such particle versus k. Give the approximate expression relating N to µ and V (r). I.4. Deduce that the volume density of fermions is given by n(r) = n0 exp(−βV (r))

,

where β = 1/kB T and n0 is a density that will be expressed as an integral. I.5. Show that at very small potential energy, such that βV (r) 1 for any r, then n0 = N/Ω. II. From now on, we consider an illuminated semiconductor, of volume Ω, maintained at temperature T . It contains N electrons, of charge −e, in its conduction band together with P = N holes (missing electrons) in its valence band. The holes will be taken as fermions, of spin 1/2, mass mh and charge +e. We assume that the high temperature limit is realized. Assume that a ﬁxed point charge Q is present at the coordinates origin, which, if alone, would create an electrostatic potential : Φext (r) =

Q 4πε0 εr r

where εr is the relative dielectric constant of the considered medium. In fact, the external charge Q induces inhomogeneous charge densities −en(r) and +ep(r) inside both fermion gases. These induced charges in turn generate electrostatic potentials, which modify the densities of mobile charges, and so on. You are going to solve this problem by a self-consistent method in the following way : let Φtot (r) be the total electrostatic potential, acting on the particles.

234

Exercises and Problems

II.1. Express the potential energies, associated with Φtot (r), to which the electrons and the holes are respectively subjected. Deduce from I. the expressions n(r) and p(r) of the corresponding densities of mobile particles. II.2. The electrostatic potential Φtot (r) is a solution of the Poisson equation, in which the charge density includes all the charges of the problem. Admit that the charge density Qδ(r) is associated with the point charge Q located at r, where δ(r) is the Dirac delta distribution. Show that, in the high temperature limit and to the lowest order in Φtot (r), the total electrostatic potential is solution of ∆Φtot (r) =

Φtot (r) Qδ(r) − λ2 ε0 εr

where λ is a constant that will be expressed versus N/Ω and T . What is the unit of λ ? II.3. Estimate λ for N/Ω = 1022 m−3 , T = 300 K, εr = 12.4. II.4. The solution of this equation is Φtot (r) =

Q exp(−r/λ) 4πε0 εr r

Indicate the physical implications of this result in a few sentences.

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” I. Consider an assembly of N magnetic ions at temperature T inside a volume Ω. These ions are independent and distinguishable. The electronic structure of an ion consists in a fundamental non-degenerate level |0, of energy ε0 , and a doubly degenerate excited state of energy ε0 + λ. Under the eﬀect of a static and uniform magnetic ﬁeld B, the fundamental level remains at the energy ε0 while the excited level splits into two sublevels |+ and |− of respective energies ε0 + λ + γB and ε0 + λ − γB. I.1. Without any calculation, plot the location of the states |0, |+ and |− versus the applied magnetic ﬁeld, in the following special cases : a) λ = 0 , kB T γB b) λ kB T γB c) kB T λ γB d) γB λ kB T

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” Conductor

235

In each case indicate which states are occupied at temperature T . I.2. Calculate the system free energy, assuming that the only degrees of freedom are those, of electronic origin, related to the three levels |0, |+ and |−. ∂F and Deduce the algebraic expressions of the system magnetization M = − ∂B M of its susceptibility χ = limB→0 (Ω is the volume). ΩB I.3. On the expressions of the magnetization and of the susceptibility obtained γB = in I.2. only discuss the limit cases a), b), c), d) of question I.1 (take kB T x). Verify that the obtained results are in agreement with the qualitative considerations of I.1.

Susceptibility

I.4. The ﬁgure below represents the variation versus temperature of the paramagnetic susceptibility of the organic “quasi-one-dimensional” compound [HN(C2 H5 )3 ](TCNQ)2 . Using the results of the present exercise can you qualitatively interpret the shape of this curve ?

Temperature (K)

Fig. 1: Magnetic susceptibility of the quasi-one-dimensional organic compound versus temperature. II. Let a paramagnetic ion have a total spin S. You will admit that, in presence the levels of this ion take the energies of a static and uniform magnetic ﬁeld B, εl = lγB, where −S ≤ l ≤ S, with l = −S, (−S + 1), (−S + 2), . . . + S. Consider an ion at thermal equilibrium at the given temperature T . Calculate

236

Exercises and Problems

its free energy f and its average magnetic moment m = −

∂f . ∂B

Write the expressions of the magnetization in the following limit cases : high and low magnetic ﬁelds ; high and low temperatures. In each situation give a brief physical comment.

Exercise 2002 : Entropies of the HC Molecule 1. Let a three-dimensional ideal gas be made of N monoatomic molecules of mass m, conﬁned inside the volume Ω and at temperature T . Without demonstration recall the expressions of its entropy S3D and of its chemical potential µ3D . 2. For a two-dimensional ideal gas of N molecules, mobile on a surface A, of the same chemical species as in 1., give the expressions, versus A, m and T , of the partition function Z2D , the entropy S2D and the chemical potential µ2D . To obtain these physical parameters, follow the same approach as in the course. Which of these parameters determines whether a molecule from the ideal gas of question 1. will spontaneously adsorb onto surface A ? 3. Let us now assume that each of the N molecules from the gas is adsorbed on surface A. What is, with respect to the gas phase, the variation per mole of translation entropy if the molecules now constitute a mobile adsorbed ﬁlm on A ? 4. Numerically calculate the value of S3D (translation entropy) for one mole of HC, of molar mass M = 36.5, at 300K under the atmospheric pressure. Calculate S2D (translation entropy on a plane) for one mole of such a gas, adsorbed on a porous silicon surface : assume that, due to the surface corrugation, the average distance between molecules on the surface is equal to 1/30 of the distance in the gas phase considered above. 5. In fact, in addition to its translation degrees of freedom, the HC molecule possesses rotation degrees of freedom. Write the general expression of the partition function of N molecules of this gas versus ztr and zrot , the translation and rotation partition functions for a single molecule, accounting for the fact that the molecules are indistinguishable. The expressions of ztr and zrot will be calculated in the next questions. 6. In this question one considers a molecule constrained to rotate in a plane around a ﬁxed axis. Its inertia moment with respect to this axis is IA . The

Problem 2001 : Quantum Boxes and Optoelectronics

237

rotation angle φ varies from 0 to 2π, the classical momentum associated with φ is pφ . The hamiltonian for one molecule corresponding to this motion is given by p2φ h= 2IA The diﬀerential elementary volume of the corresponding phase space is dφ dpφ /h, where h is the Planck constant. Calculate the partition function z1rot . 7. In the case of the linear HC molecule, of inertia moment I, free to rotate in space, the rotation hamiltonian for a single molecule is written, in spherical coordinates, p2φ p2 h= θ + 2I 2I sin2 θ Here θ is the polar angle (0 ≤ θ ≤ π), φ the azimuthal angle (0 ≤ φ ≤ 2π) ; pθ (respectively pφ ) is the momentum associated with θ (respectively φ). The volume element of the phase space is now equal to dφ dpφ dθ dpθ /h2 . Calculate the corresponding partition function z2rot . Show that the dependence in temperature and in inertia moment obtained for z2rot is the expected one, when this expression of the rotation energy is taken into account. Verify that the expression found for z2rot is in agreement with the one given in the course for a linear molecule in the classical approximation (be careful, the HC molecule is not symmetrical). 8. Numerically calculate the rotation entropy by mole at 300K of a HC mole. Take I = 2.7 × 10−40 g · cm 2 .

Problem 2001 : Quantum Boxes and Optoelectronics For more clarity, comments are written in italics and questions in standard characters. Semiconducting quantum boxes are much studied nowadays, because of their possible applications in optoelectronics. The aim of the present problem is to study the rate of spontaneous emission and the radiative eﬃciency in quantum boxes : these are two important physical parameters for the elaboration of emission devices (quantum–boxes lasers). We will particularly focus on the experimental dependence of these parameters.

238

Exercises and Problems

Fig. 1 : Electron microscope image of quantum boxes of InAs in GaAs.

Quantum boxes are made of three-dimensional semiconducting inclusions of nanometer size (see Fig.1). In the studied case, the inclusions consist of InAs and the surrounding material of GaAs. The InAs quantum boxes conﬁne the electrons to a nanometer scale in three dimensions and possess a discrete set of electronic valence and conduction states ; they are often called “artiﬁcial atoms.” In the following we will adopt a simpliﬁed description of their electronic structure, as presented below in Fig. 2.

Fig. 2: Simpliﬁed model of the electronic structure of an InAs quantum box and of the ﬁlling of its states at thermal equilibrium at T = 0. We consider that the energy levels in the conduction band, labeled by a positive integer n, are equally spaced and separated by ωc ; the energy of the fundamental level (n = 0) is noted Ec . The nth level is 2(n + 1) times degenerate, where the factor 2 accounts for the spin degeneracy and (n + 1) for the orbital degeneracy. Thus a conduction state is identiﬁed through the set of three in-

Problem 2001 : Quantum Boxes and Optoelectronics

239

1 teger quantum numbers |n, l, s, with n > 0 and 0 ≤ l ≤ n, s = ± . The ﬁrst 2 valence states are described in a similar way, with analogous notations (Ev , ωv , n , l , s ).

I. Quantum Box in Thermal Equilibrium At zero temperature and thermal equilibrium, the valence states are occupied by electrons, whereas the conduction states are empty. We are going to show in this section that the occupation factor of the conduction states of the quantum box in equilibrium remains small for temperatures up to 300K. Therefore, only in this part, the electronic structure of the quantum box will be roughly described by just considering the conduction and valence band levels lying nearest to the band gap. Fig.3 illustrates the assumed electronic conﬁguration of the quantum box at zero temperature.

Fig. 3 : Electronic conﬁguration at T = 0. I.1. Give the average number of electrons in each of these two (spindegenerate) levels versus β = 1/kB T , where T is the temperature, and the chemical potential µ . I.2. Find an implicit relation deﬁning µ versus β, by expressing the conservation of the total number of electrons in the thermal excitation process. I.3. Show that, in the framework of this model, µ is equal to (Ec + Ev )/2 independently of temperature. I.4. Estimate the average number of electrons in the conduction level at 300K and conclude. For the numerical result one will take Ec − Ev = 1 eV. In optoelectronic devices, like electroluminescent diodes or lasers, electrons are injected into the conduction band and holes into the valence band of the semiconducting active medium : the system is then out of equilibrium. The return toward equilibrium takes place through electron transitions from the conduction band toward the valence band, which may be associated with light emission.

240

Exercises and Problems

Although the box is not in equilibrium, you will use results obtained in the course (for systems in equilibrium) to study the distribution of electrons in the conduction band, and that of holes in the valence band. Consider that, inside a given band, the carriers are in equilibrium among themselves, with a speciﬁc chemical potential for each band. This is justiﬁed by the fact that the phenomena responsible for transitions between conduction states, or between valence states, are much faster than the recombination of electronhole pairs. One now assumes in sections II exactly contains one electron-hole the total number of electrons contained equal to 1 ; the same assumption is valid in the valence states.

and III that the quantum box pair (regime of weak excitation) : in all the conduction states is thus for the total number of holes present

II. Statistical Study of the Occupation Factor of the Conduction Band States II.1. Write the Fermi distribution for a given electronic state |n, l, s versus β = 1/kB T and µ, the chemical potential of the conduction electrons. Show that this distribution factor fn,l,s only depends of the index n ; it will be written fn in the following. II.2. Expressing that the total number of electrons is 1, ﬁnd a relation between µ and β. II.3. Show that µ < Ec . II.4. Show that fn is a decreasing function of n. Deduce that fn ≤

1 (n + 1)(n + 2)

(Hint : give an upper bound and a lower bound for the population of electronic states of energy smaller or equal to that of level n). II.5. Deduce that for states other than the fundamental one, it is very reasonable to approximate the Fermi-Dirac distribution function by a MaxwellBoltzmann distribution and this independently of temperature. Deduce that the chemical potential is deﬁned by the following implicit relation : 1 = g(β, w) , with 2

w = eβ(µ−Ec )

and g(β, w) =

w +w (n + 1)e−βnωc w+1 n≥1

(1)

Problem 2001 : Quantum Boxes and Optoelectronics

241

II.6. Show that the function g deﬁned in II.5 is an increasing function of w and a decreasing function of β. Deduce that the system fugacity w decreases when the temperature increases and that µ is a decreasing function of T . II.7. Specify the chemical potential limits at high and low temperatures. II.8. Deduce that beyond a critical temperature Tc – to be estimated only in the next questions – it becomes legitimate to make a Maxwell-Boltzmann-type approximation for the occupation factor of the fundamental level. II.9. Now one tries to estimate Tc . Show that expression (1) can also be written under the form : w w 1 = −w+ 2 w+1 (1 − e−βωc )2 [Hint : as a preliminary, calculate the sum χ(β) =

(2) e−βnωc and compare

n≥1

it to the sum entering into equation (1)]. II.10. Consider that the Maxwell-Boltzmann approximation is valid for the 1 fundamental state if f0 ≤ . Hence deduce Tc versus ωc . e+1 II.11. For a temperature higher than Tc show that : fn = e−βnωc

(1 − e−βωc )2 2

(3)

III. Statistics of holes III.1. Give the average number of electrons in a valence band state |n , l , s of energy En , versus the chemical potential µv of the valence electrons. (Recall : µv is diﬀerent from µ because the quantum box is out of equilibrium when an electron-hole pair is injected into it.) III.2. Deduce the average number of holes hn ,l ,s present in this valence state of energy En . Show that the holes follow a Fermi-Dirac statistics, if one associates the energy −En to this ﬁctitious particle. Write the hole chemical potential versus µv . It will thus be possible to easily adapt the results of section II, obtained for the conduction electrons, to the statistical study of the holes in the valence band.

242

Exercises and Problems

IV. Radiative lifetime of the conﬁned carriers Consider a quantum box containing an electron-hole pair at the instant t = 0 ; several physical phenomena can produce its return to equilibrium. In the present section IV we only consider the radiative recombination of the charge carriers through spontaneous emission : the deexcitation of the electron from the conduction band to the valence band generates the emission of a photon. However an optical transition between a conduction state |n, l, s and a valence state |n , l , s is only allowed if speciﬁc selection rules are satisﬁed. For the InAs quantum boxes, one can consider, to a very good approximation, that for an allowed transition n = n , l = l and s = s and that each allowed optical transition has the same strength. Consequently, the probability per unit time of radiative recombination for the electron-hole pair is simply given by fn,l,s · hn,l,s 1 = τr τ0

(4)

n,l,s

The time τr is called the radiative lifetime of the electron-hole pair ; the time τ0 is a constant characteristic of the strength of the allowed optical transitions.

nsec

IV.1. Show that, if only radiative recombinations take place and if the box contains one electron-hole pair at t = 0, the average number of electrons in the conduction band can be written e−t/τr for t > 0.

Fig. 4: Lifetime τ and radiative yield η measured for InAs quantum boxes ; the curves are the result of the theoretical model developed in the present problem. The scales of τ, η and T are logarithmic. IV.2. Now the temperature is zero. Calculate the radiative lifetime using (4). The lifetime τ of the electron-hole pair can be measured. Its experimental variation versus T is shown in Fig. 4.

Problem 2001 : Quantum Boxes and Optoelectronics

243

IV.3. Assume that at low temperature one can identify the lifetime τ and the radiative lifetime τr . Deduce τ0 from the data of Fig. 4. IV.4. From general arguments show that the radiative lifetime τr is longer for T = 0 than for T = 0. (Do not try to calculate τr in this question.) Now the temperature is in the 100–300K range. For InAs quantum boxes, some studies show that the energy spacing ω between quantum levels is of the order of 15 meV for the valence band and of 100 meV for the conduction band. IV.5. Estimate Tc for the conduction electrons and the valence holes using the result of II.10. Show that a zero temperature approximation is valid at room temperature for the distribution function of the conduction states. What is the approximation then valid for the valence states ? IV.6. Deduce the dependence of the radiative lifetime versus temperature T for T > Tc . IV.7. Compare to the experimental result for τ shown in Fig. 4. Which variation is explained by our model ? What part of the variation remains unexplained ? IV.8. From the experimental values of τ at T = 200 K and at T = 0 K deduce an estimate of the energy spacing between hole levels. Compare to the value of 15 meV suggested in IV.4.

V. Non-Radiative Recombinations

trap trap

Fig. 5 : A simple two-step mechanism of non-radiative recombination. The electron-hole pair can also recombine without emission of a photon (nonradiative recombination). A simple two-step mechanism of non-radiative recombination is schematized as follows : (a) a trap located in the neighborhood of the quantum box captures the electron ; (b) the electron then recombines with

244

Exercises and Problems

the hole, without emission of a photon. The symmetrical mechanism relying on the initial capture of a hole is also possible. One expects that such a phenomenon will be the more likely as the captured electron (or hole) lies in an excited state of the quantum box, because of its more extended wave function of larger amplitude at the trap location. The radiative and non-radiative recombinations are mutually independent. Write 1/τnr for the probability per unit time of non-radiative recombination of the electron-hole pair ; the total recombination probability per unit time is equal to 1/τ . The dependence of the emission quantum yield η of the InAs quantum boxes is plotted versus T in the Fig. 4 above : η is deﬁned as the fraction of recombinations occuring with emission of a photon. V.1. Consider a quantum box containing an electron-hole pair at the instant t = 0. Write the diﬀerential equation describing the evolution versus time of the average number of electrons in the conduction band, assuming that the electron-hole pair can recombine either with or without emission of radiation. V.2. Deduce the expressions of τ and η versus τr and τnr . V.3. Comment on the experimental dependence of η versus T . Why is it consistent with the experimental dependence of τ ? Assume that the carriers, electrons or holes, can be trapped inside the nonradiative recombination center only if they lie in a level of the quantum box of large enough energy (index n higher than or equal to n0 ). Write the nonradiative recombination rate under the following form : 1 =γ τnr

(fn,l,s + tn,l,s )

(5)

n≥n0 ,l,s

V.4. From the experimental results deduce an estimate of the index n0 for the InAs quantum boxes. [Hint : work in a temperature range where the radiative recombination rate is negligible with respect to the non-radiative recombination rate and write a simple expression versus 1/T , to ﬁrst order in η, then in ln η. You can assume, and justify the validity of this approximation, that the degeneracy of any level of index higher than or equal to n0 can be replaced by 2(n0 + 1)]. This simple description of the radiative and non-radiative recombination processes provides a very satisfactory understanding of the experimental dependences of the electron-hole pairs lifetime and of the radiative yield of the InAs

Problem 2002 : Physical Foundations of Spintronics

245

quantum boxes (see Fig. 4). These experimental results and their modeling are taken from a study performed at the laboratory of Photonics and Nanostructures of the CNRS (French National Center of Scientiﬁc Research) at Bagneux, France.

Problem 2002 : Physical Foundations of Spintronics A promising research ﬁeld in material physics consists in using the electron spin observable as a medium of information (spintronics). This observable has the advantage of not being directly aﬀected by the electrostatic ﬂuctuations, contrarly to the space observables (position, momentum) of the electron. Thus, whereas in a standard semiconductor the momentum relaxation time of an electron does not exceed 10−12 sec, its spin relaxation time can reach 10−8 sec. Here we will limit ourselves to the two-dimensional motion of electrons (quantum well structure). The electron eﬀective mass is m. The accessible domain is the rectangle [0, Lx ] × [0, Ly ] of the xOy plane. The periodic limit conditions will be chosen at the edge of this rectangle. Part II can be treated if the results indicated in part I are assumed.

Part I : Quantum Mechanics

Consider the motion of an electron described by the hamiltonian ˆ2 ˆ 0 = p + α (ˆ σx pˆx + σ H ˆy pˆy ) 2m

α≥0

is the momentum operator of the electron and the σ where pˆ = −i∇ ˆi (i = x, y, z) represent the Pauli matrices. Do not try to justify this hamiltonian. Assume that at any time t the electron state is factorized into a space part and a spin part, the space part being a plane wave of wave vector k. This normalized state is thus written

eik·r eik·r [a+ (t)|+ + a− (t)|−] = Lx Ly Lx Ly

Å

ã a+ (t) a− (t)

with |a+ (t)|2 + |a− (t)|2 = 1

(1) (2)

where |± are the eigenvectors of σ ˆz . One will write kx = k cos φ, ky = k sin φ, with k ≥ 0, 0 ≤ φ < 2π.

246

Exercises and Problems

I.1. Brieﬂy justify that one can indeed search a solution of the Schroedinger equation under the form given in (2), where k and φ are time-independent. Write the evolution equations of the coeﬃcients a± (t). I.2

ã a+ (t) des(a) For ﬁxed k and φ, show that the evolution of vector a (t) Å− ã a+ (t) d = cribing the spin state can be put under the form : i dt a− (t) ã Å a+ (t) ˆs ˆ s is a 2 × 2 hermitian matrix, that will be expliH where H a− (t) citly written versus k, φ, and the constants α, and m. ˆ s are : E± = (b) Show that the two eigenenergies of the spin hamiltonian H 2 k 2 ± αk with the corresponding eigenstates : 2m ã ã Å Å 1 1 1 1 √ |χ (3) = |χ+ = √ − eiφ −eiφ 2 2 Å

(c) Deduce that the set of |Ψk,± states :

eik·r |Ψk,+ = 2LxLy

Å

1 eiφ

ã

eik·r |Ψk,− = 2Lx Ly

Å

1 −eiφ

ã (4)

with ki = 2πni /Li (i = x, y and ni positive, negative or null integers) constitute a basis, well adapted to the problem, of the space of the one-electron states. I.3 Assume that at t = 0 the state is of the type (2) and write ω = αk. (a) Decompose this initial state on the |Ψk,± basis. (b) Deduce the expression of a± (t) versus a± (0). I.4 One deﬁnes sz (t), the average value at time t of the z-component of the ˆz . Show that : electron spin Sˆz = (/2) σ

sz (t) = sz (0) cos(2ωt) + sin(2ωt) Im a+ (0)∗ a− (0) e−iφ lying along the z-axis. The I.5 The system is installed in a magnetic ﬁeld B on the orbital variables is neglected. The hamiltonian thus becomes eﬀect of B ˆ =H ˆ 0 − γB Sˆz . H (a) Why does the momentum k remain a good quantum number ? Deduce ˆ s(B) can still be deﬁned. You that, for a ﬁxed k, a spin hamiltonian H ˆ s , γ, B and σ will express it versus H ˆz . (B) ˆ s are (b) Show that the eigenenergies of H (B)

E±

=

» 2 k 2 ± α2 k 2 + γ 2 B 2 /4 2m

Problem 2002 : Physical Foundations of Spintronics

247

(γB αk). The expression of the eigenvectors (c) Consider a weak ﬁeld B ˆ of Hs to ﬁrst order in B are then given by (B)

|χ+ = |χ+ −

γB |χ− 4αk

(B)

|χ− = |χ− +

γB |χ+ 4αk

(B) Deduce that the average values of Sˆz in the states |χ± are (B)

(B)

χ+ |Sˆz |χ+ = −

γB 4αk

(B)

(B)

χ− |Sˆz |χ− = +

(B)

γB 4αk

(B)

(B)

(d) What are the values of |χ± and of the matrix elements χ± |Sˆz |χ± , (such that γB αk). in the case of a strong ﬁeld B Part II : Statistical Physics In this part, we are dealing with the properties of an assembly of N electrons, assumed to be without any mutual interaction, each of them being submitted to the hamiltonian studied in Part I. We will ﬁrst assume that there is no magnetic ﬁeld ; the energy levels are then distributed into two branches as obtained in I.2 : E± (k) =

2 k 2 ± αk 2m

(5)

with k = |k|. We will write k0 = mα/ and 0 = 2 k02 /(2m) = mα2 /2. II.1. We ﬁrst consider the case α = 0. Brieﬂy recall why the density of states D0 () of either branch is independent of for this two-dimensional problem. Show that D0 () = mLx Ly /(2π2 ). II.2. We return to the real problem with α > 0 and now consider the branch E+ (k). (a) Plot the dispersion law E+ (k) versus k. What is the energy range accessible in this branch ? (b) Express k versus k0 , 0 , m, and the energy for this branch. Deduce the density of states : ã Å … 0 (6) D+ () = D0 () 1 − + 0 Plot D+ ()/D0 () versus /0 . II.3. One is now interested by the branch E− (k). (a) Plot E− (k) versus k and indicate the accessible energy range. How many k’s are possible for a given energy ?

248

Exercises and Problems

(b) Show that the density of states of this branch is given by … 0 < 0 : D− () = 2D0 () + 0 Å ã … 0 > 0 : D− () = D0 () 1 + + 0

(7)

II.4. In the rest of the problem, the temperature is taken as zero. (a) Qualitatively explain how the Fermi energy changes with the electrons number N . Show that if N is small enough, only the branch E− is ﬁlled. (b) What is the value of the Fermi energy F when the E+ branch begins to be ﬁlled ? Calculate the number of electrons N ∗ at this point, versus Lx , Ly , m, and 0 . (c) Where are the wave vectors k corresponding to the ﬁlled states F < 0 located ? II.5. Assume in what follows that F > 0. (a) What are the wave vectors k for the occupied states in either branch E± ? (b) Calculate the Fermi wave vectors kF+ and kF− versus F , k0 , 0 , m and . (c) For N > N ∗ , calculate N − N ∗ and deduce the relation : Lx Ly m (20 + F ) π2 II.6. A weak magnetic ﬁeld is now applied on the sample. N=

(a) Using the result of I.5, explain why the Fermi energy does not vary to ﬁrst order in B. (b) Calculate the magnetization along z of the electrons gas, to ﬁrst order in B. II.7. Now one analyzes the possibility to maintain a magnetization in the system for some time in the absence of magnetic ﬁeld. At time t = 0, each electron is prepared in the spin state |+z . The magnetization along z is thus M (0) = N /2. For a given electron, the evolution of sz (t) in the absence of collisions was calculated in I.4. Now this evolution is modiﬁed due to the eﬀect of collisions. Because of the elastic collisions of an electron on the material defects, its k wave vector direction may be randomly modiﬁed. These collisions are modeled assuming that : – they occur at regular times : t1 = τ , t2 = 2τ ,..., tn = nτ ,... – the angle φn characterizing the k direction between the nth and the (n + 1)th collision is uniformly distributed over the range [0, 2π[ and decorrelated from the previous angles φn−1 , φn−2 , ..., and thus from the coeﬃcients a± (tn ). – the angles φn corresponding to diﬀerent particles are not correlated.

Problem 2002 : Physical Foundations of Spintronics

249

Under these assumptions, the total magnetization after many collisions can be calculated by making an average over angles for each electron. One writes s¯(t) for the average of sz (t) over the angles φn . Show that s¯(tn+1 ) = cos(2ωτ ) s¯(tn ). II.8. One assumes that ωτ 1. (a) Is there a large modiﬁcation of the average spin s¯ between two collisions ? (b) One considers large times t with respect to the interval τ between two collisions. Show that the average spin s¯(t) exponentially decreases and express the characteristic decrease time td versus ω and τ . (c) Take two materials : the ﬁrst one is a good conductor (a few collisions take place per unit time), the second one is a poor conductor (many collisions occur per unit time). In which of these materials is the average spin better “protected,” i.e., has a slower decrease ? Comment on the result. (d) For τ = 10−12 sec and ω=5 µeV, estimate the spin relaxation time td . II.9. To maintain a stationary magnetization in the population of N spins, one injects g± electrons per unit time, of respective spins sz = ±/2 into the system. For each spin value, the electrons leave the system with a probability per unit time equal to 1/tr . Besides, one admits that the relaxation mechanism studied in the previous question is equivalent to the exchange between the + and − spin populations : a + spin is transformed into a − spin with a probability per unit time equal to 1/td , and conversely a − spin is transformed into a + spin with a probability per unit time equal to 1/td. One calls f± (t) the populations of each of the spin states at time t, td is the spin relaxation time. ∗ (a) Write the rate equations on df± /dt. What are the spin populations f± in steady state ? ∗ f ∗ − f− (b) Deduce the relative spin population, that is, the quantity P = + ∗ ∗. f+ + f− In what type of material will the magnetization degrade not too fast ?

Solution of the Exercises and Problems

Electrostatic Screening : exercise 2000 Magnetic Susceptibility of a “Quasi-1D” Conductor : exercise 2001 Entropies of the HC Molecule : exercise 2002 Quantum Boxes and Optoelectronics : problem 2001 Physical Fundations of Spintronics : problem 2002

251

Exercise 2000 : Electrostatic Screening

253

Exercise 2000 : Electrostatic Screening I.1. The density N/Ω and the chemical potential µ are related by N 1 = Ω Ω

1 D(k)fF D (k)d3k = 4π 3

+∞

0

4πk 2 dk ã 2 k 2 exp β −µ +1 2me Å

I.2. In the high temperature and low density limit, the exponential in the denominator of the Fermi-Dirac distribution is large with respect to 1, so that Å 2 2 ã +∞ −β k − µ N 1 2me 4πk 2 dk e Ω 4π 3 0 Å ã N 2πme kB T 3/2 2eβµ Ω h2 I.3. The total energy in presence of a potential is written ε=

2 k 2 + V (r) 2me

At high temperature and low density

N

1 4π 3

e−βV (r) d3r

Å

+∞

4πk 2 dk e

−β

ã 2 k 2 −µ 2me

0

Ω

I.4. By deﬁnition one has

n(r)d3r

N= Ω

one thus identiﬁes n(r) to n(r) = n0 e−βV (r) with n0 =

1 4π 3

Å

+∞

4πk 2 dk e

−β

0

I.5. If βV (r) 1 for any r, n is uniform, i.e., n = n0 =

N Ω

ã 2 k 2 −µ 2me

254

Solution of the Exercises and Problems

II.1. The potential energy of an electron inside the electrostatic potential Φtot (r) is −e Φtot (r), the energy of a hole +e Φtot (r). The electron and hole densities are respectively equal to n(r) = n0 exp[−β(−e Φtot (r)] p(r) = p0 exp[−β(+e Φtot (r)] with n0 = p0 =

N . Ω

II.2. The total density of charges at r is equal to e[p(r) − n(r)]. The charge localized at the origin also enters into the Poisson equation : ∆Φtot (r) +

e[p(r) − n(r)] Qδ(r) + =0 ε0 εr ε0 εr

In the high temperature limit and to the lowest order p(r) − n(r) = n0 e−βeΦtot (r) − eβeΦtot (r) −2n0 eβΦtot (r) Whence the equation satisﬁed by Φtot (r) : ∆Φtot (r) = One writes λ2 = II.3. If

2n0 e2 β Q Φtot (r) − δ(r) ε0 εr ε0 εr

Ω ε0 εr kB T , this is the square of a length. N 2e2

N = 1022 m−3 T = 300 K, εr = 12.4, then λ = 30 nm. Ω

II.4. The eﬀect of the charge located at the origin cannot be felt beyond a distance of the order of λ : at a larger distance Φtot (r) is rapidly constant, the electron and hole densities are constant andopposite. The mobile electrons and holes have screened the charge at the origin.

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” Conductor

255

Exercise 2001 : Magnetic Susceptibility of a “Quasi-One-Dimensional” Conductor I.1. a) λ = 0 kB T γB

b) λ kB T γB |+1>

ε

ε

ε0 + λ

|+1>

|−1> |0 > B

ε0

ε0

0

|0 > B

0

|−1> Three levels in ε0 for B = 0 Comparable occupations of the three levels

At the temperature T , the upper levels are weakly occupied

c) kB T λ γB |+1> ε

ε

ε0 + λ

d) γB λ kB T |+1>

ε0 + λ |−1>

ε0

0

|0 > B

ε0

0

|0 > B |−1>

Comparable occupations of the three states at T . Only level | − 1 is populated. Little inﬂuence of λ, this case The magnetization is saturated, each is almost equivalent to a) ion carries a moment +γ

256

Solution of the Exercises and Problems

I.2. State |0, energy ε0 |+ ε0 + λ + γB |− ε0 + λ − γB For a single ion : z = e−βε0 + e−β(ε0 +λ+γB) + e−β(ε0 +λ−γB) = e−βε0 [1 + e−β(λ+γB) + e−β(λ−γB) ] For the whole system : Z = (z)N F = −kB T ln Z ¶ î ó© = −N kB T −βε0 + ln 1 + e−β(λ+γB) + e−β(λ−γB) ó î = N ε0 − N kB T ln 1 + e−β(λ+γB) + e−β(λ−γB) −βλ+βγB − e−βλ−βγB e ∂F = Nγ M =− ∂B 1 + e−β(λ+γB) + e−β(λ−γB) For B → 0, M ∼

N γ2βγBe−βλ i.e., 1 + 2e−βλ χ=

2N γ 2 βe−βλ Ω(1 + 2e−βλ )

I.3. Discussion of the limit cases : (a) λ = 0 ex − e−x with x = βγB M = Nγ 1 + ex + e−x 2N γ 2 (Curie law). In the limit x → 0, one obtains χ = 3ΩkB T (b) βλ 1 βγB = x M N γe−βλ (ex − e−x ) χ

2N γ 2 e−βλ ΩkB T

(Curie law attenuated by the thermal activation towards the magnetic level). (c) 1 βλ x M

2N γ 2 βB N γ[x − βλ + βλ + x] = 3 3 χ=

This situation is analogous to a).

2N γ 2 3ΩkB T

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” Conductor

257

(d) βγB βλ 1 M Nγ The magnetization is saturated. I.4. The high temperature behavior of χ in the ﬁgure suggests a Curie law. The expression of χ obtained in I.2. vanishes for β → ∞ and for β → 0. It is positive and goes through a maximum for β of the order of λ1 (exactly for 1 − βλ+ 2e−βλ = 0, i.e., βλ = 1.47). From the temperature of the χ maximum λ can be deduced. Note : in fact, for the system studied in the present experiment, the ε0 + λ level contains 3 states, and not 2 like in this exercise based on a simpliﬁed model. II. z=

+S

e−lβγB = eβSγB

−S

2S

e−lβγB = eβSγB

l=0

1 − e−(2S+1)βγB 1 − e−βγB

ñ

1 − e−(2S+1)x f = −kB T ln z = −kB T SβγB + ln 1 − e−x m=−

ô with x = βγB

∂f ∂f (2S + 1)γβe−(2S+1)x kB T βγe−x = − γβ = γS + kB T − ∂B ∂x 1 − e−x 1 − e−(2S+1)x 2S + 1 1 m = S + (2S+1)x − γ e − 1 ex − 1

3 limits : x → +∞ x → −∞ x → 0 (a) x → +∞ (high positive ﬁeld, low temperature) m →S γ

saturation

(b) x → −∞ (high negative ﬁeld, low temperature) m → S − (2S + 1) + 1 = −S γ

saturation

(c) x → 0 : (low ﬁeld, high temperature). One expects m to be proportional to B, which allows to deﬁne a susceptibility. Therefore one has to develop m up to the term in x, which requires to express

258

Solution of the Exercises and Problems

the denominators to third order. 2S + 1 m S+ − 2 3 γ (2S + 1)x + (2S + 1)2 x2 + (2S + 1)3 x6 + ... x + S(S + 1) γBS(S + 1) m x = γ 3 3kB T

x2 2

1 3 + x6 + . . .

(Curie law)

Exercise 2002 : Entropies of the HC Molecule ò 5 Ω 3/2 + N kB (2πmk T ) B N h3 2 ò ï Ω 3/2 = −kB T ln (2πmk T ) B N h3 ï

1.

S3D = N kB ln µ3D 2. Z2D S2D µ2D

Å

ãN A 2πmkB T h2 ï ò A = N kB ln (2πmkB T ) + 2N kB N h2 ï ò A = −kB T ln (2πmk T ) B N h2 1 = N!

A molecule is spontaneously adsorbed if µ2D < µ3D . 3. If the adsorbed ﬁlm is mobile on the surface : ∆S = S2D − S3D ï ò Ah 1 ∆S = N kB ln − N kB 2 Ω(2πmkB T )1/2 4. For the only translation degrees of HC : S3D = 152.9 J.K−1 S2D = 48.2 J.K−1 The average 3D distance is of 3.3 nm, it is 0.11 nm between molecules on the surface (which is less than an interatomic distance, but one has also to account for the large corrugation of the surface).

Problem 2001 : Quantum Boxes and Optoelectronics

5. ZN =

259

1 (zrot )N (ztr )N N!

Ç å βp2φ dφdpφ exp − z1rot = h 2IA å Ç +∞ βp2φ 2π dpφ exp − = h −∞ 2IA φ 2π = 2πIA kB T h å Ç ã Å βp2φ dθdpθ dφdpφ βp2θ exp − z2rot = exp − h2 2I 2I sin2 θ

6.

7.

=

8π 2 IkB T h2

4π 2 kB T = 2 IkB T , for a symmetrical mo2hcB h lecule, such that its physical situation is identical after a 180˚ rotation. But HC is not symmetrical The expression in the course is

8. Rotation entropy of HC : Frot = −N kB T ln z2rot (the factor for indistinguishability is included inside the translation term). ò ï ∂Frot ∂ ln z2rot = N kB ln z2rot + T Srot = − ∂T ∂T −1 Srot = 33.3 J · K

Problem 2001 : Quantum Boxes and Optoelectronics I.1. Owing to the spin degeneracies one has nEc =

2 eβ(Ec −µ)

+1

and nEv =

2 eβ(Ev −µ)

+1

260

Solution of the Exercises and Problems

I.2. At any temperature nEc + nEv = 2, in the present approximation in which the total number of electrons is 2. It follows that 1 eβ(Ec −µ)

+1

+

1 eβ(Ev −µ)

+1

= 1,

an implicit relation between µ and β

I.3. This equation can be reexpressed, noting x = eβµ , y = eβEc and z = eβEv : x x + =1 y+x z+x that is, zx + yx + 2x2 = zy + zx + yx + x2 , x2 = yz, whence the result Ec + Ev , independently of temperature. µ= 2 I.4. One immediately deduces from I.3 nEc =

2 eβ(Ec −µ)

2

=

+1

e

β(Ec −Ev ) 2

+1

2 = 4 × 10−9 . e20 + 1

At 300 K with Ec − Ev = 1 eV, nEc =

At 300 K, at thermal equilibrium, the probability for the conduction levels to be occupied is thus extremely small. II.1. fnls =

1 eβ(Enls −µ)

+1

As the energy Enls only depends of n, consequently we will write fnls = fn . II.2. Noting gn for the degeneracy of state n, one has gn fn = 2(n + 1)fn n = 1 = n

n

an implicit relation between µ and β. II.3. If one had µ > Ec , 2f0 would be larger than 1, in contradiction with n = 1. II.4. En is an increasing function of n, as is also eβ(En−µ) , whence the result. One has f0 < 1/2, and fn+1 < fn , so that : 1≥

n i=0

gi fi >

n

gi fn = 2fn

i=0

n

(n + 1) = 2fn

i=0

i.e., fn ≤

1 (n + 1)(n + 2)

(n + 1)(n + 2) 2

Problem 2001 : Quantum Boxes and Optoelectronics

261

1 1 , f2 ≤ , etc. One then 6 12 β(E1 −µ) β(E2 −µ) > 5, in f2 e > 11, which justiﬁes deduces that in f1 one has e the use of the Maxwell-Boltzmann approximation, i.e.,

II.5. The relation obtained in II.4 implies f1 ≤

fnls =

1 eβ(Enls−µ)

+1

≈ e−β(En −µ)

for any excited state. Taking into account the validity of this approximation, the implicit relation between µ and β becomes 1 = 2f0 + 2

∞

(n + 1)e−β(En−µ)

n=1

With En = Ec +nωc , and w = e−β(Ec −µ) , this expression can be transformed into ∞ 1 1 (n + 1)e−βnωc = g(β, w) = −1 +w 2 w +1 n=1

=

∞ w +w (n + 1)e−βnωc w+1 n=1

(1)

II.6. Obviously g is a decreasing function of β at ﬁxed w. Moreover, ∞ ∂g 1 = + (n + 1)e−βnωc > 0, thus g is an increasing function ∂w (w + 1)2 n=1 of w. When T increases, β decreases. For g to remain constant [equal to 1/2 from (1)], the fugacity w must decrease when T increases. 1 ∂w ∂µ = w 2 (Ec − µ) + β . From II.3., µ < E0 = Ec ∂T kT ∂T ∂w < 0 implies and the ﬁrst term of the right member is > 0. Consequently, ∂T ∂µ < 0. ∂T

Since w = e−β(Ec −µ) ,

II.7. From equation (1), – when β → 0,

∞

(n + 1)e−βnωc tends to inﬁnity, thus the solution w must

n=1

tend to 0. Then µ tends to −∞. (One ﬁnds back the classical limit at high temperature) ; w 1 = , so that w 2 w+1 tends to 1 in this limit : the low temperature limit of µ is then E0 = Ec .

– when T → 0, the sum tends to 0 and (1) becomes

262

Solution of the Exercises and Problems

1 . At low temperature, w tends to 1 and f0 tends to 1/2, w−1 + 1 whereas at high temperature w tends to 0 and the Maxwell-Boltzmann statistics is then applicable to the 0 level. Since w is a monotonous function of T , this latter limit must be valid beyond a critical temperature Tc , so that then eβ(E0 −µ) 1 (w 1). II.8. f0 =

II.9. The function χ =

∞

e−βnωc is equal to

n=1

derivative

e−βωc 1 . Its = βωc −βω c 1−e e −1

∞ ∂χ = −ωc ne−βnωc ∂β n=1

allows to conveniently calculate the sum in the right member of (1), which is equal to : ∞

(n + 1)e−βnωc = χ −

n=1

=

1 eβωc 1 ∂χ = βωc + βωc ωc ∂β e − 1 (e − 1)2

2eβωc − 1 e2βωc 1 = −1 + = −1 + βω 2 βω (e c − 1) (e c − 1)2 (1 − e−βωc )2

whence the result (2). II.10. This condition is equivalent to w < 1/e, and at T = Tc one has w = 1/e. For w = 1/e, (2) becomes 1 e(e − 1) = ωc 2 2(e + 1) 1 − exp − kB Tc ωc 2 2(e + 1) = 1 − exp − i.e., 2 e +e+2 kB Tc 1+

which provides :

ωc 2(e + 1) = − ln 1 − kTc e2 + e + 2

Numerically kB Tc = 0.65 ωc. II.11. For T > Tc , the Maxwell-Boltzmann statistics is applicable to the 1 fundamental state and consequently to any state. f0 = −1 can be apw +1 proached by w, which allows to write (2) as : w 1 = 2 (1 − e−βωc )2

Problem 2001 : Quantum Boxes and Optoelectronics

263

It follows that fn = e−β(E0 +nωc −µ) = we−βnωc = e−βnωc

(1 − e−βωc )2 2

(3)

III.1. The average number of electrons in a state of spin s of energy E in the valence band is also the Fermi-Dirac distribution function : n =

1 eβ(E−µv )

+1

III.2. hn s = 1 − n =

eβ(E−µv ) 1 1 = −β(E−µ ) = β(−E+µ ) β(E−µ ) v v v + 1 e +1 e +1 e

hn s is indeed an expression of the Fermi-Dirac statistics, under the condition of writing Et = −E. The holes chemical potential is then given by µt = −µv . The results on the electrons in questions II.1. to II.11. apply to the holes (provided this transposition for the energy and for µ). IV.1. If all recombinations are radiative, dn n =− dt τr whence the result : n = n0 e−t/τr with n0 = 1. IV.2. At T = 0, both f0 and h0 are equal to 1/2, all the other occupation factors being equal to zero. Consequently, one has : f0,0,1 h0,0,1 + f0,0,−1 h0,0,−1 2f0 h0 1 1 = = = τr τ0 τ0 2τ0

i.e., τr = 2τ0

IV.3. One deduces from the ﬁgure τr = 2τ0 = 1 nsec, that is, τ0 = 0.5 nsec. IV.4. At any temperature one has fn,l,s hn,l,s 1 = τr τ0

(4)

n,l,s

Moreover the average numbers of electrons and holes are equal to 1. As the occupation factors are spin-independent, for a given value of s one gets : n,l

fn =

n ,l

hn =

1 2

264

One deduces :

Solution of the Exercises and Problems

n,l

f n hn ≤

n,l

fn

1 hn = 4 n ,l

This limit being reached at T = 0, one then ﬁnds that at any temperature the radiative lifetime is larger than or equal to its value at T = 0. IV.5. For the electrons ωc 100 meV, whereas ωv 15 meV for the holes. Tc is thus of the order of 780 K for the electrons and 117 K for the holes. Expression (2) shows that when kB T ω, w is of the order of 1, which indeed constitutes the zero temperature approximation. The MaxwellBoltzmann approximation is valid for any hole state if T > Tc = 117 K, from II.10. IV.6. One thus has f0 = 1/2 and all the other fn ’s are equal to zero, whereas from (3), (1 − e−βωv )2 h0 = 2 One then deduces that the only active transition is associated to the levels |001 and |00 − 1 (like at zero temperature) and that 1 2f0 h0 h0 (1 − e−βωv )2 = = = τr τ0 τ0 2τ0 IV.7. This model allows one to qualitatively understand the increase of the radiative lifetime with temperature, due to the redistribution of the hole probability of presence on the excited levels. This is indeed observed above 250 K, a temperature range in which a decrease of the lifetime is observed in the experiment. IV.8. At 200 K one approximately gets τr (T ) 1 (1 − e−βωv )2 = , which gives 2 1 βωv = − ln 1 − √ 2

=

2τr (0), that is,

i.e., a splitting between holes levels of 1.23 kB T for T = 200 K, that is, of 20 meV. 1 . This τr means that the variation of the average number of electrons per unit time, due to these radiative recombinations, is given by

V.1. The probability per unit time of radiative recombination is

n dn =− dt τr

Problem 2001 : Quantum Boxes and Optoelectronics

265

In presence of non radiative recombinations at the rate becomes :

1 , the total variation τnr

dn n n =− − dt τr τnr V.2. This allows to deﬁne the total lifetime τ by dn n n n =− − =− dt τr τnr τ i.e., 1 1 1 = + τ τr τnr and a radiative yield η equal to the ratio between the number of radiative recombinations per unit time and the total number of recombinations per unit time : η=

n τr n τ

=

τ = τr

1 τr 1 τr

+

1 τnr

=

1 r 1 + ττnr

V.3. When only radiative recombinations are present, τnr → ∞ and the radiative yield η is equal to 1, independently of the value of the radiative lifetime. Here the yield strongly decreases above 200 K, whereas it is constant for T < 150 K. This implies that a nonradiative recombination mechanism becomes active for T > 200 K. Correlatively, this is consistent with the decrease of the total lifetime τ in the same temperature range, whereas the results of question IV.6. show that the radiative lifetime should increase with temperature. V.4. We found in V.2 : η =

In this expression

1 = γ τnr

1 τr 1 τr

+

1 τnr

.

(fn,l,s + hn,l,s ). At suﬃciently high tem-

n≥n0 ,l,s

perature, around 350 K, τ ≈ 0.5 nsec whereas τr > 2 nsec. In this regime, τnr the nonradiative recombinations determine τ , and η ≈ . Besides, only the τr hole levels of index n0 > 0 are occupied, they satisfy a Maxwell-Boltzmann statistics. This allows to write, using (3), (1 − e−βωv )2 1 =γ 2(n + 1)e−βnωv τnr 2 n≥n0

266

Solution of the Exercises and Problems

Taking an orbital degeneracy equal to n0 + 1 for each excited level, one has (1 − e−βωv )2 1 2(n0 + 1)e−βn0 ωv =γ e−βnωv τnr 2 n≥0

= γ(1 − e

−βωv

)(n0 + 1)e

−βn0 ωv

Using the result of IV.6, 1 γ(1 − e−βωv )(n0 + 1)e−βn0 ωv (n0 + 1)e−βn0 ωv = = 2γτ0 −βωv )2 (1−e η (1 − e−βωv ) 2τ0

≈ 2γτ0 (n0 + 1)e−βn0 ωv In this temperature range ln(η) = constant +

n0 ωv . kT

The slope of η(T ) provides an approximation for n0 : between 200 and 350 K ã Å Å ã 1 n0 ωv 1 η200K − = ln(10) = ln η350K kB 200 350 which yields a value of n0 ranging between 5 and 6.

Problem 2002 : Physical Foundations of Spintronics Part I : Quantum Mechanics

I.1. A state having a plane wave eik·r for space part is an eigenstate of the momentum operator, with the eigenvalue k. The eﬀect of the hamiltonian on this type of state gives : ñ Å ãô ï 2 2Å ã Å −iφ ãò eik·r eik·r k a+ e a+ a− ˆ H = + kα a− a− eiφ a+ Lx Ly Lx Ly 2m One thus sees that, if the coeﬃcients a± verify the evolution equations ia˙ ± =

2 k 2 a± + kα e∓iφ a∓ 2m

the proposed eigenvector is a solution of the Schroedinger equation ˙ = H|ψ. ˆ i|ψ

Problem 2002 : Physical Foundations of Spintronics

267

I.2. (a) For ﬁxed k and φ, one is restricted to a two-dimensional problem, only on the spin variables, and the hamiltonian is equal to ã Å 2 k 2 0 e−iφ ˆ Hs = Id + kα eiφ 0 2m where Id represents the identity matrix. (b) It can be immediately veriﬁed that the two proposed vectors are eigenˆ s with the eigenenergies E± . vectors of H (c) The space of the one-electron states is the tensor product of the space associated to the electron motion in the rectangle [0, Lx] × [0, Ly ] with the spin space. The functions eik·r / Lx Ly are an orthonormal basis of the orbital space, provided that the choice of k allows this function to satisfy the periodical limit conditions : this requires ki = 2πni /Li , with ni positive, negative or null integer. As for the two vectors |χ± , they constitute a basis of the spin space. Thus the vectors |Ψk,± are an ˆ 0. orthonormal eigenbasis of the hamiltonian H I.3. (a) The development of the initial spin state on the pair of eigenstates |χ± is ã Å ã Å 1 1

a+ (0) = a+ (0) + a− (0)e−iφ a− (0) eiφ 2 Å ã 1

1 + a+ (0) − a− (0)e−iφ −eiφ 2 The initial state (orbital ⊕ spin) is thus given by |Ψ =

a+ (0) + a− (0)eiφ a+ (0) − a− (0)eiφ √ √ |Ψk,+ + |Ψk,− . 2 2

(b) One deduces that the electron spin state at time t is characterized by Å

a+ (t) a− (t)

ã = +

Å ã e−iE+ t/

1 −iφ a+ (0) + a− (0)e eiφ 2 Å ã e−iE− t/

1 −iφ a+ (0) − a− (0)e −eiφ 2

After development, one ﬁnds a± (t) = e−i

2 2

k t / 2m

a± (0) cos(ωt) − ia∓ (0)e∓iφ sin(ωt) .

268

Solution of the Exercises and Problems

I.4. A simple calculation gives sz (t) = =

|a+ (t)|2 − |a− (t)|2 2

cos(2ωt) |a+ (0)|2 − |a− (0)|2 2

+2 sin(2ωt) Im a∗+ (0)a− (0)e−iφ

One then deduces the result announced in the text. I.5. (a) The new term in the hamiltonian, arising from the magnetic ﬁeld, commutes with the momentumÅ operator. One Å can thus ã ã always look for eia+ a+ i k· r genstates of the form e , where is a two-component a− a− vector providing the spin state. The hamiltonian determining the evolution of this spin is written in the |±z basis : Å ã 2 k 2 −γB/2 kαe−iφ ˆ s − γB σ ˆ s(B) = H ˆz = Id + H kαeiφ γB/2 2 2m (b) One veriﬁes that the eigenenergies of the above matrix are those given in the text. σz |χ∓ = 1. One deduces (c) One notices that χ± |Sˆz |χ± = 0 and that χ± |ˆ the formulae given in the text. ˆ s(B) are practically equal to the (d) For a large B, the eigenstates of H ˆ eigenstates |±z of Sz . More precisely, for γ < 0 (electron), one has (B) (B) (B) |χ± ±|±z . Consequently, χ± |Sˆz |χ± ±/2. Part II : Statistical Physics

II.1. For α = 0, one has E± = 2 k 2 /(2m). The number of quantum states in a given spin state, with a wave vector of coordinates in the range (kx to kx +dkx ; ky to ky + dky ) is equal to d2 N =

Lx Ly Lx Ly dkx dky = k dk dφ 2 (2π) (2π)2

i.e., by integration over the angle φ : dN =

Lx Ly Lx Ly m kdk = d 2π 2π 2

The density of states is thus independent of the energy and equal to D0 () = Lx Ly m/(2π2 ).

Problem 2002 : Physical Foundations of Spintronics

269

II.2. (a) The dispersion law in this branch is written E+ (k) =

2 (k + k0 )2 − 0 2m

This corresponds to a portion of parabola, plotted in Fig.1. The spectrum of the allowed energies is [0, ∞[.

Dispersion relations E

E+

0

E

k1

k F+

k2

k0

k

F

Fig. 1 : Typical behavior of the dispersion laws E± (k) in either branch.

(b) For the branch = E+ (k), … k = −k0 + and thus

2m ( + 0 ) 2

ã Å … 0 Lx Ly dk D+ () = k = D0 () 1 − 2π d + 0

The corresponding curve is plotted in 2.

270

Solution of the Exercises and Problems 4

D(E)

3

2

1

0

1

0

1

2

3

4

5

6

7

8

9

10

E

Fig. 2: Densities of states D+ /D0 (lower curve) and D− /D0 (upper curve), plotted versus /0 . II.3. (a) For the branch = E− (k), E− (k) =

2 (k − k0 )2 − 0 2m

the variation of which is given in Fig.1. The spectrum of the permitted energies is [−0 , ∞]. – For −0 < ≤ 0, there are two possible wavevector moduli k for a given energy : … 2m k1,2 = k0 ± ( + 0 ) 2 – For > 0, there is a single allowed wavevector modulus : … 2m ( + 0 ) k = k0 + 2 (b) One then deduces, for > 0 Å ã … 0 D− () = D0 () 1 + + 0 and for < 0 (adding both contributions corresponding for either sign in the k1,2 formula) : … 0 D− () = 2D0 () + 0

Problem 2002 : Physical Foundations of Spintronics

271

The corresponding curve is plotted in Fig. 2. II.4. (a) Assume that the electrons are accommodated one next to the other in the available energy levels. Since the temperature is taken to be zero, the ﬁrst electron must be located on the fundamental level, the second one on the ﬁrst excited level, and so forth. Thus the ﬁrst levels to be ﬁlled are those of the E− branch, with k around k0 , the Fermi energy F being a little larger than −0 for N small. When N increases, the energy F also increases. It reaches 0 when all the states of the E− branch, of wave vectors ranging between 0 and 2k0 , are ﬁlled. For larger numbers of electrons, both branches E± are ﬁlled. (b) The E+ branch remains empty until F reaches the value F = 0. There are then 0 2Lx Ly m D− ()d = 0 N∗ = π2 −0 electrons on the E− branch. (c) In the E− branch, the ﬁlled states verify E− (k) k1 < k < k2 with : … 2m k1,2 = k0 ± (F + 0 ) 2

Advanced Texts in Physics This program of advanced texts covers a broad spectrum of topics that are of current and emerging interest in physics. Each book provides a comprehensive and yet accessible introduction to a field at the forefront of modern research. As such, these texts are intended for senior undergraduate and graduate students at the M.S. and Ph.D. levels; however, research scientists seeking an introduction to particular areas of physics will also benefit from the titles in this collection.

Claudine Hermann

Statistical Physics Including Applications to Condensed Matter

With 63 Figures

Claudine Hermann Laboratoire de Physique de la Matie` re Condense´ e Ecole Polytechnique 91128 Palaiseau France [email protected]

Library of Congress Cataloging-in-Publication Data is available.

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Table of Contents Introduction

xi

Glossary

xv

1 Statistical Description of Large Systems. Postulates 1.1 Classical or Quantum Evolution of a Particle ; Phase Space 1.1.1 Classical Evolution . . . . . . . . . . . . . . . . . . . 1.1.2 Quantum Evolution . . . . . . . . . . . . . . . . . . 1.1.3 Uncertainty Principle and Phase Space . . . . . . . . 1.1.4 Other Degrees of Freedom . . . . . . . . . . . . . . . 1.2 Classical Probability Density ; Quantum Density Operator . 1.2.1 Statistical Approach for Macroscopic Systems . . . . 1.2.2 Classical Probability Density . . . . . . . . . . . . . 1.2.3 Density Operator in Quantum Mechanics . . . . . . 1.3 Statistical Postulates ; Equiprobability . . . . . . . . . . . . 1.3.1 Microstate, Macrostate . . . . . . . . . . . . . . . . 1.3.2 Time Average and Ensemble Average . . . . . . . . 1.3.3 Equiprobability . . . . . . . . . . . . . . . . . . . . . 1.4 General Properties of the Statistical Entropy . . . . . . . . 1.4.1 The Boltzmann Deﬁnition . . . . . . . . . . . . . . 1.4.2 The Gibbs Deﬁnition . . . . . . . . . . . . . . . . . . 1.4.3 The Shannon Deﬁnition of Information . . . . . . . Summary of Chapter 1 . . . . . . . . . . . . . . . . . . . . . . . . Appendix 1.1 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 1.2 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 1.3 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 2 The 2.1 2.2 2.3

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Diﬀerent Statistical Ensembles. General Methods Energy States of an N -Particle System . . . . . . . . . . . . . . Isolated System in Equilibrium : “Microcanonical Ensemble” . Equilibrium Conditions for Two Systems in Contact . . . . . . 2.3.1 Equilibrium Condition : Equal β Parameters . . . . . . 2.3.2 Fluctuations of the Energy Around its Most Likely Value v

1 2 2 4 4 5 5 6 8 9 10 10 11 12 15 15 15 18 19 21 27 29 31 32 35 36 36 37

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2.4

Contact with a Heat Reservoir, “Canonical Ensemble” . . . . . 2.4.1 The Boltzmann Factor . . . . . . . . . . . . . . . . . . . 2.4.2 Energy, with Fixed Average Value . . . . . . . . . . . . 2.4.3 Partition Function Z . . . . . . . . . . . . . . . . . . . . 2.4.4 Entropy in the Canonical Ensemble . . . . . . . . . . . 2.4.5 Partition Function of a Set of Two Independent Systems 2.5 Grand Canonical Ensemble . . . . . . . . . . . . . . . . . . . . 2.5.1 Equilibrium Condition : Equality of Both T ’s and µ’s . 2.5.2 Heat Reservoir and Particles Reservoir . . . . . . . . . . 2.5.3 Grand Canonical Probability and Partition Function . . 2.5.4 Average Values . . . . . . . . . . . . . . . . . . . . . . . 2.5.5 Grand Canonical Entropy . . . . . . . . . . . . . . . . . 2.6 Other Statistical Ensembles . . . . . . . . . . . . . . . . . . . . Summary of Chapter 2 . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 2.1 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .

40 41 42 44 45 46 48 49 50 51 52 53 53 55 57

3 Thermodynamics and Statistical Physics 3.1 Zeroth Law of Thermodynamics . . . . . . . . . . . . . . . . . . 3.2 First Law of Thermodynamics . . . . . . . . . . . . . . . . . . . 3.2.1 Work . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.2.2 Heat . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 3.2.3 Quasi-Static General Process . . . . . . . . . . . . . . . 3.3 Second Law of Thermodynamics . . . . . . . . . . . . . . . . . 3.4 Third Law of Thermodynamics . . . . . . . . . . . . . . . . . . 3.5 The Thermodynamical Potentials ; the Legendre Transformation 3.5.1 Isolated System . . . . . . . . . . . . . . . . . . . . . . . 3.5.2 Fixed N , Contact with a Heat Reservoir at T . . . . . . 3.5.3 Contact with a Heat and Particle Reservoir at T . . . . 3.5.4 Transformation of Legendre ; Other Potentials . . . . . . Summary of Chapter 3 . . . . . . . . . . . . . . . . . . . . . . . . . .

59 60 60 61 63 63 64 66 67 67 69 70 72 75

4 The Ideal Gas 4.1 Introduction . . . . . . . . . . . . . . . . . . . . . . . . . 4.2 Kinetic Approach . . . . . . . . . . . . . . . . . . . . . . 4.2.1 Scattering Cross Section, Mean Free Path . . . . 4.2.2 Kinetic Calculation of the Pressure . . . . . . . . 4.3 Classical or Quantum Statistics ? . . . . . . . . . . . . . 4.4 Classical Statistics Treatment of the Ideal Gas . . . . . 4.4.1 Calculation of the Canonical Partition Function . 4.4.2 Average Energy ; Equipartition Theorem . . . . . 4.4.3 Free Energy ; Physical Parameters (P, S, µ) . . . 4.4.4 Gibbs Paradox . . . . . . . . . . . . . . . . . . . 4.5 Conclusion . . . . . . . . . . . . . . . . . . . . . . . . . Summary of Chapter 4 . . . . . . . . . . . . . . . . . . . . . . Appendix 4.1 . . . . . . . . . . . . . . . . . . . . . . . . . . .

79 79 80 80 81 83 85 85 87 89 90 91 93 95

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vii

Appendix 4.2 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 103 5 Indistinguishability, the Pauli Principle 5.1 Introduction . . . . . . . . . . . . . . . . . . . . . 5.2 States of Two Indistinguishable Particles . . . . . 5.2.1 General Case . . . . . . . . . . . . . . . . 5.2.2 Independent Particles . . . . . . . . . . . 5.3 Pauli Principle ; Spin-Statistics Connection . . . 5.3.1 Pauli Principle ; Pauli Exclusion Principle 5.3.2 Theorem of Spin-Statistics Connection . . 5.4 Case of Two Particles of Spin 1/2 . . . . . . . . . 5.4.1 Triplet and Singlet Spin States . . . . . . 5.4.2 Wave Function of Two Spin 1/2 Particles 5.5 Special Case of N Independent Particles . . . . . 5.5.1 Wave Function . . . . . . . . . . . . . . . 5.5.2 Occupation Numbers . . . . . . . . . . . . 5.6 Return to the Introduction Examples . . . . . . 5.6.1 Fermions Properties . . . . . . . . . . . . 5.6.2 Bosons Properties . . . . . . . . . . . . . Summary of Chapter 5 . . . . . . . . . . . . . . . . . .

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113 114 116 116 116 119 119 120 121 121 123 124 124 125 126 126 126 129

6 General Properties of the Quantum Statistics 131 6.1 Use of the Grand Canonical Ensemble . . . . . . . . . . . . . . 132 6.1.1 2 Indistinguishable Particles at T , Canonical Ensemble . 132 6.1.2 Description in the Grand Canonical Ensemble . . . . . . 133 6.2 Factorization of the Grand PartitionFunction . . . . . . . . . . 134 6.2.1 Fermions and Bosons . . . . . . . . . . . . . . . . . . . . 134 6.2.2 Fermions . . . . . . . . . . . . . . . . . . . . . . . . . . 135 6.2.3 Bosons . . . . . . . . . . . . . . . . . . . . . . . . . . . . 135 6.2.4 Chemical Potential and Number of Particles . . . . . . . 136 6.3 Average Occupation Number ;Grand Potential . . . . . . . . . . 136 6.4 Free Particle in a Box ; Density of States . . . . . . . . . . . . . 138 6.4.1 Quantum States of a Free Particle in a Box . . . . . . . 138 6.4.2 Density of States . . . . . . . . . . . . . . . . . . . . . . 143 6.5 Fermi-Dirac Distribution ; Bose-Einstein Distribution . . . . . . 147 6.6 Average Values of Physical Parameters at T . . . . . . . . . . . 148 6.7 Common Limit of the Quantum Statistics . . . . . . . . . . . . 149 6.7.1 Chemical Potential of the Ideal Gas . . . . . . . . . . . 149 6.7.2 Grand Canonical Partition Function of the Ideal Gas . . 151 Summary of Chapter 6 . . . . . . . . . . . . . . . . . . . . . . . . . . 153 7 Free Fermions Properties 155 7.1 Properties of Fermions at Zero Temperature . . . . . . . . . . . 155 7.1.1 Fermi Distribution, Fermi Energy . . . . . . . . . . . . . 155 7.1.2 Internal Energy and Pressure at Zero Temperature . . . 158

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7.1.3 Magnetic Properties. Pauli Paramagnetism . . . . . . . Properties of Fermions at Non-Zero Temperature . . . . . . . . 7.2.1 Temperature Ranges and Chemical Potential Variation . 7.2.2 Speciﬁc Heat of Fermions . . . . . . . . . . . . . . . . . 7.2.3 Thermionic Emission . . . . . . . . . . . . . . . . . . . . Summary of Chapter 7 . . . . . . . . . . . . . . . . . . . . . . . . . . Appendix 7.1 . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 7.2

8 Elements of Bands Theory and Crystal Conductivity 8.1 What is a Solid, a Crystal ? . . . . . . . . . . . . . . . . . 8.2 The Eigenstates for the Chosen Model . . . . . . . . . . . 8.2.1 Recall : the Double Potential Well . . . . . . . . . 8.2.2 Electron on an Inﬁnite and Periodic Chain . . . . 8.2.3 Energy Bands and Bloch Functions . . . . . . . . . 8.3 The Electron States in a Crystal . . . . . . . . . . . . . . 8.3.1 Wave Packet of Bloch Waves . . . . . . . . . . . . 8.3.2 Resistance ; Mean Free Path . . . . . . . . . . . . . 8.3.3 Finite Chain, Density of States, Eﬀective Mass . . 8.4 Statistical Physics of Solids . . . . . . . . . . . . . . . . . 8.4.1 Filling of the Levels . . . . . . . . . . . . . . . . . 8.4.2 Variation of Metal Resistance versus T . . . . . . . 8.4.3 Insulators’ Conductivity Versus T ; Semiconductors 8.5 Examples of Semiconductor Devices . . . . . . . . . . . . 8.5.1 The Photocopier : Photoconductivity Properties . 8.5.2 The Solar Cell : an Illuminated p − n Junction . . 8.5.3 CD Readers : the Semiconductor Quantum Wells . Summary of Chapter 8 . . . . . . . . . . . . . . . . . . . . . . .

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159 160 160 163 165 169 171 175 176 177 177 180 182 184 184 184 185 189 189 190 191 196 196 196 197 199

9 Bosons : Helium 4, Photons,Thermal Radiation 201 9.1 Material Particles . . . . . . . . . . . . . . . . . . . . . . . . . . 201 9.1.1 Thermodynamics of the Boson Gas . . . . . . . . . . . . 202 9.1.2 Bose-Einstein Condensation . . . . . . . . . . . . . . . . 203 9.2 Bose-Einstein Distribution of Photons . . . . . . . . . . . . . . 206 9.2.1 Description of the Thermal Radiation ; the Photons . . 206 9.2.2 Statistics of Photons, Bosons in Non-Conserved Number 207 9.2.3 Black Body Deﬁnition and Spectrum . . . . . . . . . . . 210 9.2.4 Microscopic Interpretation . . . . . . . . . . . . . . . . . 212 9.2.5 Photometric Measurements : Deﬁnitions . . . . . . . . . 214 9.2.6 Radiative Balances . . . . . . . . . . . . . . . . . . . . . 217 9.2.7 Greenhouse Eﬀect . . . . . . . . . . . . . . . . . . . . . 218 Summary of Chapter 9 . . . . . . . . . . . . . . . . . . . . . . . . . . 221 Solving Exercises and Problems of Statistical Physics

223

Units and Physical Constants

225

Table of Contents

ix

A few useful formulae

229

Exercises and Problems Ex. 2000 : Electrostatic Screening . . . . . . . . . . Ex. 2001 : Magnetic Susceptibility of a “Quasi-1D” Ex. 2002 : Entropies of the HC Molecule . . . . . Pr. 2001 : Quantum Boxes and Optoelectronics . . Pr. 2002 : Physical Foundations of Spintronics . .

. . . . . . . Conductor . . . . . . . . . . . . . . . . . . . . .

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231 233 234 236 237 245

Solution of the Exercises and Problems Ex. 2000 : Electrostatic Screening . . . . . . . . . . Ex. 2001 : Magnetic Susceptibility of a “Quasi-1D” Ex. 2002 : Entropies of the HC Molecule . . . . . Pr. 2001 : Quantum Boxes and Optoelectronics . . Pr. 2002 : Physical Foundations of Spintronics . .

. . . . . . . Conductor . . . . . . . . . . . . . . . . . . . . .

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251 253 255 258 259 266

Index

275

Introduction A glosssary at the end of this introduction deﬁnes the terms speciﬁc to Statistical Physics. In the text, these terms are marked by an asterisk in exponent. A course of Quantum Mechanics, like the one taught at Ecole Polytechnique, is devoted to the description of the state of an individual particle, or possibly of a few ones. Conversely, the topic of this book will be the study of systems ∗ containing very many particles, of the order of the Avogadro number N , for example the molecules in a gas, the components of a chemical reaction, the adsorption sites for a gas on a surface, the electrons of a solid. You certainly previously studied this type of system, using Thermodynamics which is ruled by “exact” laws, such as the ideal gas one. Its physical parameters, that can be measured in experiments, are macroscopic quantities like its pressure, volume, temperature, magnetization, etc. It is now well-known that the correct microscopic description of the state of a system, or of its evolution, requires the Quantum Mechanics approach and the solution of the Schroedinger equation, but how can this equation be solved when such a huge number of particles comes into play ? Printing on a listing the positions and velocities of the N molecules of a gas would take a time much longer that the one elapsed since the Big Bang ! A statistical description is the only issue, which is the more justiﬁed as the studied system is larger, since the relative ﬂuctuations are then very small (Ch. 1). This course is restricted to systems in thermal equilibrium ∗ : to reach such an equilibrium it is mandatory that interaction terms should be present in the hamiltonian of the total system : even if they are weak, they allow energy exchanges between the system and its environment (for example one may be concerned by the electrons of a solid, the environment being the ions of the same solid). These interactions provide the way to equilibrium. This approach takes place during a characteristic time, the so-called “relaxation time”, with a range of values which depends on the considered system (the study of oﬀequilibrium phenomena, in particular transport phenomena, is another branch

xi

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Introduction

of the ﬁeld of Statistical Physics, which is not discussed in this book). Whether the system is in equilibrium or not, the parameters accessible to an experiment are just those of Thermodynamics. The purpose of Statistical Physics is to bridge the gap between the microscopic modeling of the system and the macroscopic physical parameters that characterize it (Ch. 2). The studied system can be found in a great many states, solutions of the Quantum Mechanics problem, which diﬀer by the values of their microscopic parameters while satisfying the same speciﬁed macroscopic physical conditions, for example a ﬁxed number of particles and a given temperature. The aim is thus to propose statistical hypotheses on the likehood for a particular microscopic state or another one to be indeed realized in these conditions : the statistical description of systems in equilibrium is based on the postulate on the quantity statistical entropy : it should be maximum consistently with the constraints on the system under study. The treated problems concern all kinds of degrees of freedom : translation, rotation, magnetization, sites occupied by the particles, etc. For example, for a system made of N spins, of given total magnetization M , from combinatory arguments one will look for all the microscopic spin conﬁgurations leading to M ; then one will make the basic hypothesis that, in the absence of additional information, all the possible conﬁgurations are equally likely (Ch. 2). Obviously it will be necessary to verify the validity of the microscopic model, as this approach must be consistent with the laws and results of Thermodynamics (Ch. 3) It is speciﬁed in the Quantum Mechanics courses that the limit of Classical Mechanics is justiﬁed when the de Broglie wavelength associated with the wavefunction is much shorter than all the characteristic dimensions of the problem. When dealing with free indistinguishable mobile particles, the characteristic length to be considered is the average distance between particles, which is thus compared to the de Broglie wavelength, or to the size of a typical wave packet, at the considered temperature. According to this criterion, among the systems including a very large number of mobile particles, all described by Quantum Mechanics, two types will thus be distinguished, and this will lead to consequences on their Statistical Physics properties : – in dilute enough systems the wave packets associated with two neighboring particles do not overlap. In such systems, the possible potential energy of a particle expresses an external ﬁeld force (for example gravity) or its conﬁnement in a ﬁnite volume. This system constitutes THE “ideal gas” : this means that all diluted systems of mobile particles, subjected to the same external potential, have the same properties, except for those related to the particle mass (Ch. 4) ;

Introduction

xiii

– in the opposite case of a high density of particles, like the atoms of a liquid or the electrons of a solid, the wave packets of neighboring particles do overlap. Quantum Mechanics analyzes this latter situation through the Pauli principle : you certainly learnt a special case of it, the Pauli exclusion principle, which applies to the ﬁlling of atomic levels by electrons in Chemistry. The general expression of the Pauli principle (Ch. 5) speciﬁes the conditions on the symmetry of the wavefunction for N identical particles. There are only two possibilities and to each of them is associated a type of particle : - on the one hand, the fermions, such that only a single fermion can be in a given quantum state ; - on the other hand, the bosons, which can be in unlimited number in a determined quantum state. The consequences of the Pauli principle on the statistical treatment of noninteracting indistinguishable particles, i.e., the two types of Quantum Statistics, that of Fermi-Dirac and that of Bose-Einstein, are ﬁrst presented in very general terms in Ch. 6. In the second part of this course (Ch. 7 and the following chapters), examples of systems following Quantum Statistics are treated in detail. They are very important for the physics and technology of today. Some properties of fermions are presented using the example of electrons in metallic (Ch. 7) or, more generally, crystalline (Ch. 8) solids. Massive boson particles, in conserved number, are studied using the examples of the superﬂuid helium and the Bose-Einstein condensation of atoms. Finally, the thermal radiation, an example of a system of bosons in non-conserved number, will introduce us into very practical current problems (Ch. 9). A topic in physics can only be fully understood after suﬃcient practice. A selection of exercises and problems with their solution is presented at the end of the book. This introductory course of Statistical Physics emphasizes the microscopic interpretation of results obtained in the framework of Thermodynamics and illustrates its approach, as much as possible, through practical examples : thus the Quantum Statistics will provide an opportunity to understand what is an insulator, a metal, a semiconductor, or what is the principle of the greenhouse eﬀect that could deeply inﬂuence our life on earth (and particularly that of our descendants !). The content of this book is inﬂuenced by the previous courses of Statistical Physics taught at Ecole Polytechnique : the one by Roger Balian From microphysics to macrophysics : methods and application to statistical physics, volume I translated by D. ter Haar and J.F. Gregg, volume II translated by D.

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Introduction

ter Haar, Springer Verlag Berlin (1991), given during the 1980s ; the course by Edouard Brézin during the 1990s. I thank them here for all that they brought to me in the stimulating ﬁeld of Statistical Physics. Several discussions in the present book are inspired from the course by F. Reif Fundamentals of Statistical and Thermal Physics, Mac Graw-Hill (1965), a not so recent work, but with a clear and practical approach, that should suit students attracted by the “physical” aspect of arguments. Many other Physical Statistics text books, of introductory or advanced level, are edited by Springer. This one oﬀers the point of view of a top French scientiﬁc Grande Ecole on the subject. The present course is the result of a collective work of the Physics Department of Ecole Polytechnique. I thank the colleagues with whom I worked the past years, L. Auvray, G. Bastard, C. Bachas, B. Duplantier, A. Georges, T. Jolicoeur, M. Mézard, D. Quéré, J-C. Tolédano, and particularly my coworkers from the course of Statistical Physics “A”, F. Albenque, I. Antoniadis, U. Bockelmann, J.-M. Gérard, C. Kopper, J.-Y. Marzin, who brought their suggestions to this book and with whom it is a real pleasure to teach. Finally, I would like to thank M. Digot, M. Maguer and D. Toustou, from the Printing Oﬃce of Ecole Polytechnique, for their expert and good-humored help in the preparation of the book.

Glossary We begin by recalling some deﬁnitions in Thermodynamics that will be very useful in the following. For convenience, we will also list the main deﬁnitions of Statistical Physics, introduced in the next chapters of this book. This section is much inspired by chapter 1, The Language of Thermodynamics, from the book Thermodynamique, by J.-P. Faroux and J. Renault (Dunod, 1997). Some deﬁnitions of Thermodynamics : The system is the object under study ; it can be of microscopic or macroscopic size. It is distinguished from the rest of the Universe, called the surroundings. A system is isolated if it does not exchange anything (in particular energy, particles) with its surroundings. The parameters (or state variables) are independent quantities which deﬁne the macroscopic state of the system : their nature can be mechanical (pressure, volume), electrical (charge, potential), thermal (temperature, entropy), etc. If these parameters take the same value at any point of the system, the system is homogeneous. The external parameters are independent parameters (volume, electrical or magnetic ﬁeld) which can be controlled from the outside and imposed to the system, to the experimental accuracy. The internal parameters cannot be controlled and may ﬂuctuate ; this is the case for example of the local repartition of density under an external constraint. The internal parameters adjust themselves under the eﬀect of a modiﬁcation of the external parameters. A system is in equilibrium when all its internal variables remain constant in time and, in the case of a system which is not isolated, if it has no exchange with its surroundings : that is, there is no exchange of energy, of electric charges, of particles. In Thermodynamics it is assumed that any system, submitted to constant and uniform external conditions, evolves toward an equilibrium state that it can no longer spontaneously leave afterward. The thermal equilibrium between two systems is realized after exchanges between themselves : this is not possible if the walls which separate them are adiabatical, i.e., they do not transmit any energy. xv

xvi

Glossary

In a homogeneous system, the intensive parameters, such as its temperature, pressure, the diﬀerence in electrical potential, do not vary when the system volume increases. On the other hand, the extensive parameters such as the volume, the internal energy, the electrical charge are proportional to the volume. Any evolution of the system from one state to another one is called a process or transformation. An inﬁnitesimal process corresponds to an inﬁnitely small variation of the external parameters between the initial state and the ﬁnal state of the system. A reversible transformation takes place through a continuous set of equilibrium intermediate states, for both the system and the surroundings, i.e., all the parameters deﬁning the system state vary continuously : it is then possible to vary these parameters in the reverse direction and return to the initial state. A transformation which does not obey this deﬁnition is said to be irreversible. In a quasi-static transformation, at any time the system is in internal equilibrium and its internal parameters are continuously deﬁned. Contrarly to a reversible process, this does not imply anything about the surroundings but only means that the process is slow enough with respect to the characteristic relaxation time of the system. Some deﬁnitions of Statistical Physics : A conﬁguration deﬁned by the data of the microscopic physical parameters, given by Quantum or Classical Mechanics, is a microstate. A conﬁguration deﬁned by the data of the macroscopic physical parameters, given by Thermodynamics, is a macrostate. An ensemble average is performed at a given time on an assembly of systems of the same type, prepared in the same macroscopic conditions. In the microcanonical ensemble, each of these systems is isolated and its energy E is ﬁxed, that is, it is lying in the range between E and E + δE. It contains a ﬁxed number N of particles. In the canonical ensemble, each system is in thermal contact with a large system, a heat reservoir, which imposes its temperature T ; the energy of each system is diﬀerent but for macroscopic systems the average E is deﬁned with very good accuracy. Each system contains N particles. In the canonical ensemble the partition function is the quantity that norms the probabilities, which are Boltzmann factors. (see Ch. 2) In the grand canonical ensemble, each system is in thermal contact with a heat

Glossary

xvii

reservoir which imposes its temperature T , the energy of each system being diﬀerent. The energies of the various systems are spread around an average value, deﬁned with high accuracy in the case of a macroscopic system. Each system is also in contact with a particle reservoir, which imposes its chemical potential : the number of particles diﬀers according to the system, the average value N is deﬁned with high accuracy for a macroscopic system.

Chapter 1

Statistical Description of Large Systems. Postulates of Statistical Physics The aim of Statistical Physics is to bridge the gap between the microscopic and macroscopic worlds. Its ﬁrst step consists in stating hypotheses about the microscopic behavior of the particles of a macroscopic system, i.e., with characteristic dimensions very large with respect to atomic distances ; the objective is then the prediction of macroscopic properties, which can be measured in experiments. The system under study may be a gas, a solid, etc., i.e., of physical or chemical or biological nature, and the measurements may deal with thermal, electrical, magnetic, chemical, properties. In the present chapter we ﬁrst choose a microscopic description, either through Classical or Quantum Mechanics, of an individual particle and its degrees of freedom. The phase space is introduced, in which the time evolution of such a particle is described by a trajectory : in the quantum case, this trajectory is deﬁned with a limited resolution because of the Heisenberg uncertainty principle (§ 1.1). Such a description is then generalized to the case of the very many particles in a macroscopic system : the complexity arising from the large number of particles will be suggested from the example of molecules in a gas (§ 1.2). For so large numbers, only a statistical description can be considered and § 1.3 presents the basic postulate of Statistical Physics and the concept of statistical entropy, as introduced by Ludwig Boltzmann (1844-1906), an Austrian physicist, at the end of the 19th century.

1

2

1.1

Chapter 1. Statistical Description of Large Systems

Classical or Quantum Evolution of a Particle ; Phase Space

For a single particle we describe the time evolution ﬁrst in a classical framework, then in a quantum description. In both cases, this evolution is conveniently described in the phase space.

1.1.1

Classical Evolution px

t1 t0

O

x

Fig. 1.1 : Trajectory of a particle in the phase space. Consider a single classical particle, of momentum p0 , located at the coordinate r0 at time t0 . It is submitted to a force F0 (t). Its time evolution can be predicted through the Fundamental Principle of Dynamics, since d p = F (t) dt

(1.1)

This evolution is deterministic since the set (r, p) can be deduced at any later time t. One introduces the one-particle “phase space”, at six dimensions, of coordinates (x, y, z, px , py , pz ). The data (r, p) correspond to a given point of this space, the time evolution of the particle deﬁnes a trajectory, schematized on the Fig. 1.1 in the case of a one-dimension space motion. In the particular case of a periodic motion, this trajectory is closed since after a period the particle returns at the same position with the same momentum : for example the abscissa x and momentum px of a one-dimension harmonic oscillator of

Classical or Quantum Evolution of a Particle

3

mass m and frequency ω are linked by : 1 p2x + mω 2 x2 = E (1.2) 2m 2 In the phase space this relation is the equation of an ellipse, a closed trajectory periodically described (Fig. 1.2). Another way of determining the time px

t0 t1 O

E

x

Fig. 1.2: Trajectory in the phase space for a one-dimension harmonic oscillator of energy E. evolution of a particle is to use the Hamilton equations (see for example the course of Quantum Mechanics by J.-L. Basdevant and J. Dalibard, including a CDROM, Springer, 2002) : the motion is deduced from the position r and its corresponding momentum p. The hamiltonian function associated with the total energy of the particle of mass m is given by : p2 + V (r) (1.3) 2m where the ﬁrst term is the particle kinetic energy and V is its potential energy, from which the force in (1.1) is derived. The Hamilton equations of motion are given by : ⎧ ∂h ⎪ ⎨ r˙ = ∂ p (1.4) ⎪ ⎩ p˙ = − ∂h ∂r and are equivalent to the Fundamental Relation of Dynamics (1.1). h=

Note : Statistical Physics also applies to relativistic particles which have a diﬀerent energy expression.

4

1.1.2

Chapter 1. Statistical Description of Large Systems

Quantum Evolution

It is well known that the correct description of a particle and of its evolution requires Quantum Mechanics and that to the classical hamiltonian function h ˆ The Schroedinger equation corresponds the quantum hamiltonian operator h. provides the time evolution of the state |ψ : i

∂|ψ ˆ = h|ψ ∂t

(1.5)

The squared modulus of the spatial wave function associated with |ψ gives the probability of the location of the particle at any position and any time. Both in Classical and Quantum Mechanics, it is possible to reverse the time direction, i.e., to replace t by −t in the motion equations (1.1), (1.4) or (1.5), and thus obtain an equally acceptable solution.

1.1.3

Uncertainty Principle and Phase Space

You know that Quantum Mechanics introduces a probabilistic character, even when the particle is in a well-deﬁned state : if the particle is not in an eiˆ the measurement result is uncertain. genstate of the measured observable A, Indeed, for a single measurement the result is any of the eigenvalues aα of the considered observable. On the other hand, when this measurement is reproduced a large number of times on identical, similarly prepared systems, the ˆ average of the results is ψ|A|ψ, a well-deﬁned value. px

h

O

x

Fig. 1.3: A phase space cell, of area h for a one-dimension motion, corresponds to one quantum state.

Classical Probability Density ; Quantum Density Operator

5

In the phase space, this quantum particle has no well-deﬁned trajectory since, at a given time t, in space and momentum coordinates there is no longer an exact position, but a typical extent of the wave function inside which this particle can be found. Another way of expressing this diﬃculty is to state the Heisenberg uncertainty principle : if the particle abscissa x is known to ∆x, then the corresponding component px of its momentum cannot be determined with an accuracy better than ∆px , such that ∆x · ∆px ≥ /2, where = h/2π, h being the Planck constant. The “accuracy” on the phase space trajectory is thus limited, in other words this trajectory is “blurred” : the phase space is therefore divided into cells of area of the order of h for a one-dimension motion, h3 if the problem is in three dimensions, and the state of the considered particle cannot be deﬁned better than within such a cell (Fig. 1.3). This produces a discretization of the phase space. (In the case of a classical particle, it is not a priori obvious that such a division is necessary, and if it is the case, that the same cell area should be taken. It will be stated in § 1.2.2 that, for sake of consistency at the classical limit of the quantum treatment, this very area should be chosen.)

1.1.4

Other Degrees of Freedom

Up to now we have only considered the degrees of freedom related to the particle position, which take continuous values in Classical Mechanics and are described by the spatial part of the wave function in Quantum Mechanics. If the particle carries a magnetic moment, the wave function includes an additional component, with discrete eigenvalues (see a Quantum Mechanics course).

1.2

Classical Probability Density ; Quantum Density Operator

When considering a macroscopic system, owing to the extremely large number of parameters, the description of the particles’ evolution can only be done through a statistical approach. The tools are diﬀerent for either the classical or the quantum description of the particles’ motion.

6

1.2.1

Chapter 1. Statistical Description of Large Systems

Necessity of a Statistical Approach for Macroscopic Systems

The physical parameters referring to individual particles, the evolution of which was described in the above section, are measured at the atomic scale (distances of the order of a Bohr radius of the hydrogen atom, i.e., a tenth of a nm, energies of the order of a Rydberg or of an electron-volt, characteristic times in the range of the time between two collisions in a gas, that is about 10−10 sec for N2 in standard conditions). For an object studied in Thermodynamics or measured in usual laboratory conditions, the scales are quite diﬀerent : the dimensions are currently of the order of a mm or a cm, the energies are measured in joules, the measurement time is of the order of a second. Thus two very diﬀerent ranges are evidenced : the microscopic one, described by Classical or Quantum Mechanics laws, unchanged when reversing the direction of time ; the macroscopic range, which is often the domain of irreversible phenomena, as you learnt in previous courses. Now we will see what are the implications when shifting from the microscopic scale to the macroscopic one. For that we will ﬁrst consider the case of a few particles, then of a number of particles of the order of the Avogadro number. If the physical system under study contains only a few particles, for example an atom with several electrons, the behavior of each particle follows the Mechanics laws given above. Once the position and momentum of each particle are known at t = t0 , together with the hamiltonian determining the system evolution, in principle one can deduce r and p at any further time t (yet the calculations, which are the task of Quantum Chemistry in particular, quickly become very complex). Thus in principle the problem can be solved, but one is facing practical diﬃculties, which quickly become “impossibilities”, as soon as the system under study contains “many” particles. Indeed in Molecular Dynamics, a branch of Physics, the evolution of particle assemblies is calculated according to the Classical Mechanics laws, using extremely powerful computers. However, considering the computing capacities available today, the size of the sample is restricted to several thousands of particles at most. Consider now a macroscopic situation, like those described by Thermodynamics : take a monoatomic gas of N molecules, where N = 6.02 × 1023 is the Avogadro number, and assume that at time t0 the set of data (r0 , p0 ) is known for each particle. Without prejudging the time required for their calculation, just printing the coordinates of all these molecules at a later time t, with a printer delivering the coordinates of one molecule per second, would take 2 × 1016 years ! The information issued in one day would be for 105 molecules only. Another way to understand the enormity of such an information on N molecules is to wonder how many volumes would be needed to print it.

Classical Probability Density ; Quantum Density Operator

7

The answer is : many more than the total of books manufactured since the discovery of printing ! One thus understands the impossibility of knowing the position and momentum of each particle at any time. In fact, the measurable physical quantities (for example volume, pressure, temperature, magnetization) only rarely concern the properties of an individual molecule at a precise time : if the recent achievements of near-ﬁeld optical microscopy or atomic-force microscopy indeed allow the observation of a single molecule, and only in very speciﬁc and favorable cases, the duration of such experiments is on a “human” scale (in the range of a second). One almost always deals with physical quantities concerning an assembly of molecules (or particles) during the observation time. “It is suﬃcient for the farmer to be certain that a cloud burst and watered the ground and of no use to know the way each distinct drop fell. Take another example : everybody understands the meaning of the word “granite”, even if the shape, chemical composition of the diﬀerent crystallites, their composition ratios and their colors are not exactly known. We thus always use concepts which deal with the behavior of large scale phenomena without considering the isolated processes at the corpuscular scale.” (W. Heisenberg, “Nature in the contemporary physical science, ” from the French translation, p.42 Idées NRF, Paris (1962)). Besides, some of the observed macroscopic properties may have no microscopic equivalent : “Gibbs was the ﬁrst one to introduce a physical concept which can apply to a natural object only if our knowledge of this object is incomplete. For example if the motions and positions of all the molecules in a gas were known, speaking of the temperature of this gas would no longer retain a meaning. The temperature concept can only be used when a system is insuﬃciently determined and statistical conclusions are to be deduced from this incomplete knowledge.” (W. Heisenberg, same reference, p.45). It is impossible to collect the detailed information on all the molecules of a gas, anyway all these data would not really be useful. A more global knowledge yielding the relevant physical parameters is all that matters : a statistical approach is thus justiﬁed from a physical point of view. Statistics are directly related to probabilities, i.e., a large number of experiments have to be repeated either in time or on similar systems (§ 1.2.2). Consequently, one will have to consider an assembly (or an “ensemble”) consisting of a large number of systems, prepared in a similar way, in which the microscopic structures diﬀer and the measured values of the macroscopic parameters are not necessarily identical : the probability of occurrence of a particular “event” or a particular

8

Chapter 1. Statistical Description of Large Systems

measurement is equal to the fraction of the ensemble systems for which it occurs (see § 1.3.2).

1.2.2

Classical Probability Density

The hamiltonian is now a function of the positions and momenta of the N i , . . . p N ). particles, i.e., H(r1 , . . . ri , . . . rN , p1 , . . . p Using a description analogous to that of the one-particle phase space of § 1.1.1, the system of N classical particles now has a trajectory in a new phase space, at 6N dimensions since there are 3N space and 3N momenta coordinates. As it is impossible to exactly know the trajectory, owing to the lack of information on the coordinates of the individual particles (see § 1.2.1), one can only speak of the probability of ﬁnding the system particles at time t in the neighborhood of a given point of the phase space, of coordinates N d3ri d3 pi . (r1 , . . . ri , . . . rN , p1 , . . . pi , . . . pN ) to i=1

On this phase space, a measure dτ and a probability density D have to be deﬁned. Then the probability will be given by D(r1 , . . . , ri , . . . , p1 , . . . , pi , . . . t)dτ . From the Liouville theorem (demonstrated in Appendix 1.1) the phase space N elementary volume d3ri d3 pi is conserved for any time t during the pari=1

ticles evolution. A consequence is that for a macroscopic equilibrium, the only situation considered in the present book, the probability density is timeindependent. The elementary volume of the N -particle phase space is homogeneous to an action to the power 3N , that is, [mass × (length)2 × (time)−1 ]3N , and thus has the same dimension as h3N , the Planck constant to the same power. The inﬁnitesimal volume will always be chosen large with respect to the volume of the elementary cell of the N -particle phase space, so that it will contain a large number of such cells. (The volume of the elementary cell will be chosen as h3N , from arguments similar to the ones for a single particle.) Finally, dτ may depend on N , the particle number, through a constant CN . Here we will take dτ =

N CN 3 3 d ri d pi . The quantity dτ is thus dimensionh3N i=1

less and is related to the “number of states” inside the volume d3ri d3 pi ; the constant CN is determined in § 4.4.3 and 6.7 so that dτ should indeed give

Classical Probability Density ; Quantum Density Operator

9

the number of states in this volume : this choice has a interpretation in the quantum description, and the continuity between classical and quantum descriptions is achieved through this choice of CN . The classical probability density D(r1 , . . . ri , . . . , p1 , . . . , pi , . . . , t) is then deﬁned : it is such that D(r1 , . . . ri , . . . , p1 , . . . , pi , . . . , t) ≥ 0, and D(r1 , . . . ri , . . . , p1 , . . . pi , . . . , t)dτ = 1 (1.6) on the whole phase space. The average value at a given time t of a physical parameter A(ri , p i , t) is then calculated using : A(t) = A(ri , p i , t)D(ri , p i , t)dτ (1.7) In the same way, the standard deviation σA , representing the ﬂuctuation of the parameter A around its average value at a given t, is calculated from : 2 (σA (t))2 = A2 (t) − A(t)2 = A(ri , p i , t) − A(ri , p i , t) D(ri , p i , t)dτ (1.8)

1.2.3

Density Operator in Quantum Mechanics

By analogy with the argument of § 1.1.2, if at time t0 the total system is in the state |ψ, now with N particles, the Schroedinger equation allows one to deduce its evolution through : i

∂|ψ ˆ = H|ψ ∂t

(1.9)

ˆ is the hamiltonian of the total system with N particles. where H It was already pointed out that, when the system is in a certain state |ψ, the result of a single measurement of the observable Aˆ on this system is uncertain ˆ : the result is one of the eigenvalues (except when |ψ is an eigenvector of A) ˆ with the probability |ψ|ϕα |2 , associated with the projection of |ψ aα of A, on the eigenvector |ϕα corresponding to the eigenvalue aα (that we take here to be nondegenerate, for sake of simplicity). If the measurement is repeated ˆ many times, the average of the results is ψ|A|ψ. Now another origin of uncertainty, of a diﬀerent nature, must also be included if the state of the system at time t is already uncertain. Assume that the

10

Chapter 1. Statistical Description of Large Systems

macroscopic system may be found in diﬀerent orthonormal states |ψn , with probabilities pn (0 ≤ pn ≤ 1). In such a case the results average for repeated measurements will be : ˆ n = ˆ n = Tr (D ˆ A) ˆ ˆ = pn ψn |A|ψ pn ψn |ψn ψn |A|ψ (1.10) A n

n

The density operator, deﬁned by ˆ = D

pn |ψn ψn |

(1.11)

n

has been introduced (its properties are given in Appendix 1.2). This operator includes both our imperfect knowledge of the system’s quantum state, through the pn factors, and the ﬂuctuations related to the quantum measurement, which would already be present if the state |ψn was certain (this state appears through its projector |ψn ψn |). One veriﬁes that ˆ n = A ˆ ˆ A) ˆ = pn ψn |ψn ψn |A|ψ (1.12) Tr (D n

n

since ψn |ψn = δn n . In the current situations described in the present course, the projection states ˆ of the N -particle total |ψn will be the eigenstates of the hamiltonian H system, the eigenvalues of which are the accessible energies of the system. In ˆ will thus be diagonal on such a basis : the problems considered the operator D each energy eigenstate |ψn will simply be weighted by its probability pn . Now hypotheses are required to obtain an expression of D(ri , pi , t) in the classical description, or of the probabilities pn in the quantum one.

1.3

Statistical Postulates ; Equiprobability

We should now be convinced that a detailed microscopic treatment is impossible for a system with a very large number of particles : we will have to limit ourselves to a statistical approach. It is necessary to deﬁne the microscopic and macroscopic states to which this description will apply.

1.3.1

Microstate, Macrostate

A conﬁguration deﬁned by the data of microscopic physical parameters (for example, positions and momenta of all the particles ; quantum numbers characterizing the state of each particle ; magnetization of each paramagnetic

Statistical Hypotheses ; Equiprobability

11

atom localized in a solid) is a microstate ∗ .1 A conﬁguration deﬁned by the value of macroscopic physical parameters (total energy, pressure, temperature, total magnetization, etc.) is a macrostate ∗ . Obviously a macrostate is almost always realized by a very large number of microstates : for example, in a paramagnetic solid, permuting the values of the magnetic moments of two ﬁxed atoms creates a new microstate associated with the same macrostate. Assume that each magnetic moment can only take two values : +µ and −µ. The number of microstates W (p) all corresponding to the macrostate = (2p − N ) of magnetization M µ, in which p magnetic moments µ are in a given direction and N − p in the opposite direction, is the number of choices of the p moments to be reversed among the total of N . It is thus equal to N! p p . Since N is macroscopic and in general p too, CN = is a very CN p!(N − p)! large number. The scope of this book is the analysis of systems in equilibrium ∗ : this means that one has waited long enough so that macroscopic physical parameters now keep the same values in time within very small ﬂuctuations (the macrostate is ﬁxed), whereas microscopic quantities may continue their evolution. Besides, even when the system is not isolated ∗ from its surroundings, if it has reached equilibrium, by deﬁnition there is no algebraical transport out of the system of matter, energy, etc. The approach toward equilibrium is achieved through interactions, bringing very small amounts of energy into play, yet essential to thermalization : in a solid, for example, electrons reach an equilibrium, among themselves and with the ions, through the collisions they suﬀer, for example, on ions vibrating at non-vanishing temperature, or impurities. Here the characteristic time is very short, of the order of 10−14 sec in copper at room temperature. For other processes this time may be much longer : the time for a drink to reach its equilibrium temperature in a refrigerator is of the order of an hour. If the system has no measurable evolution during the experiment, it is considered to be in equilibrium.

1.3.2

Time Average and Ensemble Average

Once equilibrium is reached, the time average of a physical parameter, which is likely to ﬂuctuate in a macrostate, is calculated using the classical probability density (1.7) or the quantum density operator (1.11), that is, for example

1 The

words marked by an asterisk are deﬁned in the glossary preceding Chapter 1.

12

Chapter 1. Statistical Description of Large Systems

1 A = T

t +T

dt t

A(ri , p i , t)D(ri , p i , t)dτ,

(1.13)

Would the same result be obtained by calculating the average at a given time on an assembly of macrostates prepared in a similar way ? A special case, allowing this issue to be understood, is that of an isolated system, of constant total energy E. In the N -particle phase space, the condition of constant energy deﬁnes a 6N − 1-dimension surface. The points representing the various systems prepared at the same energy E are on this surface. Is the whole surface described ? What is the time required for that ? This equivalence of the time average with the average on an ensemble of systems prepared in the same way is the object of the ergodic theorem, almost always applicable and that we will assume to be valid in all the cases treated in this course. In Statistical Physics, the common practice since Josiah Willard Gibbs (18391903) is indeed to take the average, at a ﬁxed time, on an assembly (an ensemble) of systems of the same nature, prepared under the same macroscopic conditions : this is the so-called ensemble average. As an example, consider a system, replicated a large number of times, which is exchanging energy with a reservoir. The energy distribution between the system and the reservoir differs according to the replica whereas, as will be shown in § 2.3.1, the system temperature is always that of the heat reservoir. Let N be the number of prepared identical systems, where N is a very large number. We consider a physical parameter f that can ﬂuctuate, in the present case the energy of each system. Among the N systems, this parameter takes the value fl in Nl of them ; consequently, the ensemble average of the parameter f will be equal to 1 fN = Nl fl (1.14) N l

When N tends to inﬁnity, Nl /N tends to the probability pl to achieve the value fl and the ensemble average tends to the average value calculated using these probabilities : lim fN = f = pl f l (1.15) N →∞

1.3.3

l

Equiprobability

On daily experience ( !), when throwing a die, in the absence of any other information, it is commonly assumed that every face is equally likely to occur, so that the probability of occurrence of a given face (1 for example) is 1/6.

Statistical Hypotheses ; Equiprobability

13

In Statistical Physics, an analogous postulate is stated : one assumes that in the absence of additional information, for an isolated system of energy ranging between E and E + δE, where δE is the uncertainty, all the accessible microstates are equally probable. (In a classical description, this is equivalent to saying that all the accessible phase space cells are equiprobable.) Considering the example of the magnetic moments, this means that, to obtain the same magnetization value, it is equally likely to have a localized magnetic moment reversed on a particular site or on another one. The larger the number of accessible states, the larger the uncertainty or disorder, the smaller the information about the system : a measure of the disorder is the number of accessible states, or the size of the accessible volume, of the phase space.

Ω/2

Ω/2

Fig. 1.4 : Joule-Gay-Lussac expansion. Consider for example a container of volume Ω, split into two chambers of volume Ω/2 each, isolated from its surroundings (Fig. 1.4). If the gas at equilibrium is completely located inside the left chamber, the disorder is smaller than if, at the same conditions, the equilibrium occurs in the total volume Ω. Opening the partition between the two chambers, in the so-called JouleGay-Lussac free expansion, means relaxing the constraint on the molecules to be located in the left chamber. This is an irreversible ∗ process2 as the probability for a spontaneous return, at a later time, to the situation where the N molecules all lie in the left chamber is extremely low (2−N ). It appears that a macroscopic initial state distinct from equilibrium will most probably evolve toward a larger disorder, although the ﬂuctuations in equilibrium are identical when the direction of time is reversed. The direction of time deﬁned in Statistical Physics is based on the idea that a statistically improbable state was built through an external action, requiring the expending of energy. The basic postulate of Statistical Physics, stated by Ludwig Boltzmann in 1877 for an isolated system, consists in deﬁning a macroscopic quantity from the number of accessible microstates. This is the statistical entropy S, given 2 A very interesting discussion on the character of the time direction and on irreversibility in Statistical Physics can be found in the paper by V. Ambegaokar and A.A. Clerk, American Journal of Physics, vol 67, p. 1068 (1999).

14

Chapter 1. Statistical Description of Large Systems

by S = kB ln W (E)

(1.16)

Here W (E) is the number of microstates with an energy value between E and E+δE ; the constant kB is a priori arbitrary. To make this statistical deﬁnition of the entropy coincide with its thermodynamical expression (see § 3.3), one should take kB equal to the Boltzmann constant, which is the ratio of the ideal gas constant R = 8.31 J/K to the Avogadro number N = 6.02 × 1023, that is, kB = R/N = 1.38 × 10−23J/K. This is what will be taken in this course. Take the example of the localized magnetic moments of a solid in the presence of an : in the conﬁguration where p moments are parallel external magnetic ﬁeld B and N − p moments are in the opposite direction, the magnetization is to B = (2p − N ) ·B = (N − 2p)µ · B. It was M µ and the magnetic energy E = −M N! . For a large system, shown in § 1.3.1 that in such a case W (E) = p!(N − p)! using the Stirling formula for the factorial expansion, one gets ïÅ ã Å ã E 2 kB N S = kB ln W (E) = 1+ ln 2 N µB B 1 + (E/N µB B) ã Å ãò (1.17) Å 2 E ln + 1− N µB B 1 − (E/N µB B) When two independent, or weakly coupled, subsystems are associated, with respective numbers of microstates W1 and W2 , the number of microstates of the combined system is W = W1 · W2 , since any microstate of subsystem 1 can be associated with any microstate of 2. Consequently, S = S1 + S2 , so that S is additive ∗ . A special case is obtained when the two subsystems are identical, then S = 2S1 , that is, S is extensive ∗ . Formula (1.16) is fundamental in that it bridges the microscopic world (W ) and the macroscopic one (S). It will be shown in § 3.3 that S is indeed identical to the entropy deﬁned in Thermodynamics. Before stating the general properties of the statistical entropy S as deﬁned by (1.16), we ﬁrst discuss those of W (E). The aim is to enumerate the microstates of given energy E, to the measurement accuracy δE. The number of accessible states is related to both the number of particles and to the number of degrees of freedom of each particle. It is equal to the accessible volume of the phase space, to the factor CN /h3N (§ 1.2.2). The special case of free particles, in motion in a macroscopic volume, is treated in Appendix 1.3. This example of macroscopic system allows one to understand : i) the extremely fast variation of W (E) with E, as a power of the particle number N , assumed to be very large (a similar fast variation is also obtained for an assembly of N particles subjected to a potential) ; ii) the to-

General Properties of the Statistical Entropy

15

tally negligible eﬀect of δE in practical cases ; iii) the eﬀective proportionality of ln W (E), and thus of S, to N . In the general case, the particles may have other degrees of freedom (in the example of diatomic molecules : rotation, vibration), they may be subjected to a potential energy, etc. Results i), ii), iii) are still valid.

1.4

General Properties of the Statistical Entropy

As just explained on a special case, the physical parameter that contains the statistical character of the system under study is the statistical entropy. We now deﬁne it on more general grounds and relate it to the information on the system.

1.4.1

The Boltzmann Deﬁnition

The expression (1.16) of S given by Boltzmann relies on the postulate of equal probability of the W accessible microstates of an isolated system. S increases as the logarithm of the microstates number.

1.4.2

The Gibbs Deﬁnition

A more general deﬁnition of the statistical entropy was proposed by Gibbs : S = −kB

pi ln pi

(1.18)

i

This deﬁnition applies to an assembly of systems constructed under the same macroscopic conditions, made up in a similar way according to the process deﬁned in § 1.3.2 ; in the ensemble of all the reproduced systems, pi is the probability to attain the particular state i of the system, that is, the microstate i. It is this Gibbs deﬁnition of entropy that will be used in the remaining part of this course. For example, for systems having reached thermal equilibrium at temperature T , we will see that the probability of realizing a microstate i is higher, the lower the microstate energy (pi is proportional to the Boltzmann factor, introduced in § 2.4.1). The Boltzmann expression (1.16) of the entropy is a special case, which corresponds to the situation where all W microstates are equiprobable, so that the probability of realizing each of them is pi = 1/W .

16

Chapter 1. Statistical Description of Large Systems

Then : S = −kB ln

1 ( pi ) = kB ln W W i

(1.19)

It is not possible to proceed the other way and to deduce the general Gibbs deﬁnition from the particular Boltzmann deﬁnition. Yet let us try to relate the Gibbs and the Boltzmann deﬁnitions using a particular example : we are considering a system made up of n1 + n2 equiprobable microstates regrouped into two subsystems 1 and 2. Subsystem 1 contains n1 microstates, its probability of realization is p1 = n1 /(n1 + n2 ). In the same way, subsystem 2 contains n2 microstates, its probability of realization is equal to p2 = n2 /(n1 + n2 ). From the Boltzmann formulation, the statistical entropy of the total system is Stot = kB ln(n1 + n2 )

(1.20)

This entropy is related to the absence of information on the microstate of the total system in fact realized : – we do not know whether the considered microstate belongs subsystem 1 or 2 : this corresponds to the entropy S12 , which is the Gibbs term that we want to ﬁnd in this example ; – a part S1 of the entropy expresses the uncertainty inside system 1 owing to its division into n1 microstates S1 = p1 (kB ln n1 )

(1.21)

and is weighted by the probability p1 . In the same way S2 comes from the uncertainty inside system 2 S2 = p2 (kB ln n2 )

(1.22)

Since the various probabilities of realization are multiplicative, the corresponding entropies, which are related to the probabilities through logarithms, are additive : Stot = S1 + S2 + S12 kB ln(n1 + n2 ) =

n1 n2 kB ln n1 + kB ln n2 + S12 n1 + n2 n1 + n2

(1.23)

General Properties of the Statistical Entropy

17

Then the S12 term is given by Å S12 = kB Å

n1 + n2 n1 + n2

ã ln(n1 + n2 )

− kB

ã Å ã n1 n2 ln n1 − kB ln n2 n1 + n2 n1 + n2

= −kB

n1 n2 n1 n2 ln − kB ln n1 + n2 n1 + n2 n1 + n2 n1 + n2

(1.24)

= −kB (p1 ln p1 + p2 ln p2 ) Indeed this expression is the application of the Gibbs formula to this particular case. On the other hand, expression (1.16) of the Boltzmann entropy is a special case of SGibbs , which makes (1.18) maximum : the information on the system 1 for each is the smallest when all W microstates are equiprobable and pi = W i. The Boltzmann expression is the one that brings S to its maximum with the constraint of the ﬁxed value of the system energy. The expression (1.18) of S, as proposed by Gibbs, satisﬁes the following properties : S ≥ 0, for 0 ≤ pi ≤ 1 and ln pi ≤ 0 S = 0 if the system state is certain : then p = 1 for this particular state, 0 for the other ones. S is maximum when all pi ’s are equal. S is additive : indeed, when two systems 1 and 2 are weakly coupled, so that ⎧ ⎪ ⎪ ⎨ S1

= −kB

⎪ S ⎪ ⎩ 2

= −kB

i

p1i ln p1i p2j ln p2j

(1.25)

j

by deﬁnition, the entropy of the coupled system will be written : Stot = −kB

pl ln pl

(1.26)

l

Now, since the coupling is weak, each state of the total system is characterized

18

Chapter 1. Statistical Description of Large Systems

by two particular states of subsystems 1 and 2 and pl = p1i p2j . Consequently, Stot = −kB p1i p2j (ln p1i + ln p2j ) i,j

= −kB

i

p1i ln p1i

p2j + p2j ln p2j p1i

j

j

(1.27)

i

= S1 + S 2 One thus ﬁnds the additivity of S, its extensivity being a special case that is obtained by combining two systems of the same density, but of diﬀerent volumes.

1.4.3

The Shannon Deﬁnition of Information

We just saw that the more uncertain the system state, or the larger the number of accessible microstates, the larger S ; in addition, S is zero if the system state is exactly known. The value of S is thus related to the lack of information on the system state. The quantitative deﬁnition of information, given by Claude Shannon (1949), is closely copied from the Gibbs deﬁnition of entropy.3 It analyzes the capacity of communication channels. One assumes that there are W possible distinct messages, that si (i = 0, 1, . . . W − 1) is the content of the message number i, and that the probability for si to be emitted is pi . Then the information content I per sent message is I=−

W −1

pi log2 pi

(1.28)

i=0

This deﬁnition is, to the constant kB ln2, the same as that of the Gibbs entropy, since log2 pi = lnpi /ln2. The entropy is also related to the lack of information : probabilities are attributed to the various microstates, since we do not exactly know in which of them is the studied system. These probabilities should be chosen in such a way that they do not include unjustiﬁed hypotheses, that is, only the known properties of the system are introduced, and the entropy (the missing information) is maximized with the constraints imposed by the physical conditions of the studied system. This is the approach that will be chosen in the arguments of § 2.3, 2.4, and 2.5. 3 See the paper by J. Machta in American Journal of Physics, vol 67, p.1074 (1999), which compares entropy, information, and algorithmics using the example of meteorological data.

Summary of Chapter 1 The time evolution of an individual particle is associated with a trajectory in the six-dimensional phase space, of coordinates (r, p ). The Heisenberg uncertainty principle and the classical limit of quantum properties require that a quantum state occupies a cell of area h3 in this space. At a given time t an N -particle state is represented by a point in the 6N -dimensional phase space. The classical statistical description of a macroscopic system utilizes a probability density, such that the probability of ﬁnding the system of N particles N

1 , . . . , pN ), to Π d3ri d3 pi , is in the neighborhood of the point (r1 , . . . , rN , p i=1

equal to N CN 3 1 , . . . , pN ) 3N d ri d pi D(r1 , . . . , rN , p h i=1

where CN is a constant depending on N only. In Quantum Mechanics the density operator ˆ = D pn |ψn ψn | n

is introduced, which contains the uncertainties related both to the incomplete knowledge of the system and to the quantum measurement. A conﬁguration deﬁned by the data of the microscopic physical parameters is a microstate. A macrostate is deﬁned by the value of macroscopic physical parameters ; it is generally produced by a very large number of microstates. This course will only deal with Statistical Physics in equilibrium. The time average of a ﬂuctuating physical parameter is generally equivalent to the average on an assembly of identical systems prepared in the same way (“ensemble average”). One assumes that, in the absence of additional information, in the case of an 19

20

Summary of Chapter 1

isolated system of energy between E and E + δE, where δE is the uncertainty, all the W (E) accessible microstates are equally probable. The statistical entropy S is then deﬁned by S = kB ln W (E) where kB = R/N is the Boltzmann constant, with R the ideal gas constant and N the Avogadro number. The general deﬁnition of the statistical entropy, valid for an ensemble average, is S = −kB pi ln pi i

where pi is the probability of occurrence of the particular microstate i of the system. The expression for an isolated system is a special case of the latter deﬁnition. The statistical entropy is an extensive parameter, which is equal to zero when the microscopic state of the system is perfectly known.

Appendix 1.1 The Liouville Theorem in Classical Mechanics There are two possible ways to consider the problem of the evolution between the times t and t + dt of an ensemble of systems taking part in the ensemble average, each of them containing N particles, associated with the same macrostate and prepared in the same way. To one system, in its particular microstate, corresponds one point in the 6N -dimensional phase space. Either we may be concerned with the time evolution of systems contained in a small volume around the considered point (§ A) ; or we may evaluate the number of systems that enter or leave a ﬁxed inﬁnitesimal volume of the phase space during a time interval (§ B). In both approaches it will be shown that the elementary volume of the phase space is conserved between the times t and t + dt. A. Time evolution of an ensemble of systems At time t the systems occupy the phase space volume dτ around the point with coordinates (q1 , . . . qi , . . . , q3N , p1 , . . . pi , . . . , p3N ) (Fig. 1.5). The inﬁnitesimal 3N volume has the expression dτ = CN dpi dqi , where CN is a constant that i=1

only depends on N . To compare this volume to the one at the instant t + dt, one has to evaluate 3N 3N the products dpi (t)dqi (t) and dpi (t + dt)dqi (t + dt), that is, to calculate i=1

i=1

the Jacobian of the transformation ⎧ ⎪ ⎪ pi (t) → pi (t + dt) = pi (t) + dt · ∂pi (t) ⎨ ∂t ⎪ ∂q ⎪ ⎩ qi (t) → qi (t + dt) = qi (t) + dt · i (t) ∂t 21

(1.29)

22

Appendix 1.1

pi t + dt dτ t

dτ O

qi

Fig. 1.5: The diﬀerent systems, which were located in the volume dτ of the 6N -dimensional phase space at t, are in dτ at t + dt. i.e., the determinant of the 6N × 6N matrix ⎤ ⎡ ∂qi (t + dt) ∂pi (t + dt) ⎢ ∂qj (t) ∂qj (t) ⎥ ⎥ ⎢ ⎥ ⎢ ⎣ ∂qi (t + dt) ∂pi (t + dt) ⎦ ∂pj (t) ∂pj (t)

(1.30)

The Hamilton equations are used, which relate the conjugated variables qi and pi through the N -particle hamiltonian and are the generalization of Eq. (1.4) : ⎧ ∂qi ∂H ⎪ ⎪ ⎨ ∂t = ∂p i (1.31) ⎪ ∂pi ∂H ⎪ ⎩ =− ∂t ∂qi One deduces

and in particular

⎧ ∂H ⎪ ⎪ ⎨ qi (t + dt) = qi (t) + dt · ∂p i ⎪ ∂H ⎪ ⎩ pi (t + dt) = pi (t) − dt · ∂qi

(1.32)

⎧ ∂2H ∂qi (t + dt) ⎪ ⎪ ⎨ ∂q (t) = δij + dt · ∂q ∂p j j i 2 ⎪ ∂p (t + dt) ∂ H ⎪ ⎩ i = δij − dt · ∂pj (t) ∂pj ∂qi

(1.33)

Appendix 1.1

23

The obtained determinant is developed to ﬁrst order in dt : ˆ ) = 1 + dt · Tr M ˆ + O(dt2 ) Det(1 + dt · M with ˆ =− Tr M

3N 3N ∂2H ∂2H + =0 ∂pi ∂qi i=1 ∂pi ∂qi i=1

(1.34)

Consequently, the volume around a point of the 6N -dimensional phase space is conserved during the time evolution of this point ; the number of systems in this volume, i.e., D(. . . qi , . . . , . . . pi , . . . , t)dτ , is also conserved. B. Number of systems versus time in a ﬁxed volume dτ of the phase space In the volume dτ = CN

3N

dqi · dpi , at time t, the number of systems present

i=1

is equal to D(q1 , . . . qi , . . . q3N , p1 , . . . pi , . . . p3N )dτ

(1.35)

Each system evolves over time, so that it may leave the considered volume while other systems may enter it. p1

q˙1 (t)dt dq1 dp1 (q˙1 (t) + O

∂ q˙1 (t) dq1 ) dt ∂q1 q1

Fig. 1.6: In the volume dq1 · dp1 (thick line), between the times t and t + dt, the systems which enter were in the light grey rectangle at t ; those which leave the volume were in the dark grey hatched rectangle. It is assumed that q˙1 (t) is directed toward the right. Take a “box”, like the one in Fig. 1.6 which represents the situation for the only coordinates q1 , p1 . If only this couple of conjugated variables is considered, the

24

Appendix 1.1

problem involves a single coordinate q1 in real space (so that the velocity q˙1 is necessarily normal to the “faces” at constant q1 ) and we determine the variation of the number of systems on the segment for the q1 coordinate [q1 , q1 + dq1 ]. Considering this space coordinate, between the times t and t + dt, one calculates the diﬀerence between the number of systems which enter the volume through the faces normal to dq1 and the number which leave through the same faces : the ones which enter were at time t at a distance of this face smaller or equal to q˙1 (t)dt (light grey area on the Fig.), their density is D(q1 − dq1 , . . . qi , . . . qÅ 3N , p1 , . . . pi , . . . p3Nã). Those which leave were inside the ∂ q˙1 (t) dq1 dt from the face (dark grey hatched volume, at a distance q˙1 (t) + ∂q1 area). The resulting change, between t and t + dt and for this set of faces, of the number of particles inside the considered volume, is equal to −

∂(q˙1 (t)D) dq1 dt ∂q1

(1.36)

For the couple of “faces” at constant p1 , an analogous evaluation is performed, also contributing to the variation of the number of particles in the considered “box.” Now at 3N dimensions, for all the “faces,” the net increase of the number of particles in dτ during the time interval dt is −

3N ∂(Dq˙i (t)) i=1

∂qi

dqi dt −

3N ∂(Dp˙ i (t)) i=1

∂pi

dpi dt

(1.37)

This is related to a change in the particle density in the considered volume ∂D through to dtdτ . Thus the net rate of change is given by ∂t ã ã Å 3N Å 3N ∂D ∂D ∂D ∂ q˙i ∂ p˙ i (1.38) =− q˙i + p˙ i − D + ∂t ∂qi ∂pi ∂qi ∂pi i=1 i=1 From the Hamilton equations, the factor of D is zero and the above equation is equivalent to ã 3N Å dD ∂D ∂D ∂D = + q˙i + p˙ i = 0 dt ∂t ∂qi ∂pi i=1

(1.39)

dD(q1 , . . . , qi , . . . q3N , p1 , . . . , pi , . . . p3N ) is the time derivative of the where dt probability density when one moves with the representative point of the system in the 6N -dimensional phase space. This result is equivalent to the one of § A.

Appendix 1.1

25

As an example, consider the microcanonical ensemble, in which the density D is a constant for the energies between E and E + δE, zero elsewhere. The density D is function of the sole energy, a constant in the microcanonical case, so that its derivative with respect to energy is zero between E and E + δE. Then D is independent of the coordinates qi and of the momenta pi , and from (1.39) it is time-independent.

Appendix 1.2 Properties of the Density Operator in Quantum Mechanics ˆ deﬁned by The operator D ˆ = D

pn |ψn ψn |

n

where pn is the probability to realize the microstate |ψn , has the following properties : ˆ =D ˆ+ i) it is hermitian : D Indeed, for any operator |u, ˆ u|D|u =

ˆ + |u pn |u|ψn |2 = u|D

(1.40)

n

ii) it is deﬁned positive ˆ u|D|u =

pn |u|ψn |2 ≥ 0

(1.41)

n

iii) it is normed to unity, since ˆ= Tr D

n

27

pn = 1

(1.42)

Appendix 1.3 Estimation of the Number of Microstates for Free Particles Conﬁned Within a Volume Ω Take the example of a free point particle, of energy and momentum linked p2 = ε ; in three dimensions, the particle occupies the volume Ω. The by 2m constant energy surfaces correspond to a constant p. In the 6-dimensional (r, p ) phase space for a single particle, the volume associated with the energies between ε and ε + δε is equal to : Ω4πp2 δp = Ω4πp.pδp, with … m pδp , that is, δp = δε (1.43) δε = m 2ε For a single particle, the number of accessible states between ε and ε + δε is proportional to Ωp · pδp, i.e., Ωε1/2 δε, i.e., also to Ωε(3/2)−1 δε. For N independent free particles, the constant-energy sphere of the 6N dimensional phase space contains all the states of energy smaller or equal to E, where E is of the order of N ε. The volume of such a sphere varies in ΩN p3N . The searched microstates, of energy between E and E + δE, correspond to a volume of the order of ΩN p3N −1 δp, that is, ΩN p3N −2 δE, for pδp ∝ δE. The number W (E) of microstates with energy between E and E + δE is proportional to this volume, i.e., Å W (E) =

Ω Ω0

ãN Å

E E0

ã3N/2

δE E

(1.44)

where the constants Ω0 and E0 provide the homogeneity of the expression and where δE, the uncertainty on the energy E, is very much smaller than this energy value. The increase of W (E) versus energy and volume is extremely fast : for 29

30

Appendix 1.3

example, if N is of the order of the Avogadro number, N , Å W (E) =

Ω Ω0

ã6×1023 Å

E E0

ã9×1023

δE E

(1.45)

which is such a huge number that it is diﬃcult to imagine it ! The logarithm in the deﬁnition of the Boltzmann entropy S is here : ã Å Ω E 3N δE ln W (E) = N ln ln (1.46) + + ln Ω0 2 E0 E E is of the order of N times the energy of an individual particle. For N in the N range, the dominant terms in ln W (E) are proportional to N : indeed for an experimental accuracy δE/E of the order of 10−4 , the last logarithm is equal to −9.2 and is totally negligible with respect to the two ﬁrst terms, of the order of 1023 . Note : the same type of calculation, which estimates the number of accessible states in a given energy range, will also appear when calculating the density of states for a free particle (§ 6.4).

Chapter 2

The Diﬀerent Statistical Ensembles. General Methods in Statistical Physics In the ﬁrst chapter we convinced ourselves that the macroscopic properties of a physical system can only be analyzed through a statistical approach, in the framework of Statistical Physics. Practically, the statistical description of a system with a very large number of particles in equilibrium, and the treatment of any exercise or problem of Statistical Physics, always follow the two steps (see the section General method for solving exercises and problems, at the end of Ch. 9) : – First, determination of the microscopic states (microstates) accessible by the N -particles system : this is a Quantum Mechanics problem, which may be reduced to a Classical Mechanics one in some speciﬁc cases. § 2.1 will schematically present the most frequent situations. – Second, evaluation of the probability for the system to be in a particular microstate, in the speciﬁed physical conditions : it is at this point that the statistical description comes in. To solve this type of problem, one has to express that in equilibrium the statistical entropy is maximum, under the constraints deﬁned by the physical situation under study : the system may be isolated, or in thermal contact 31

32

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

with a heat reservoir (a very large system, which will dictate its temperature), the system may exchange particles or volume with another system, and so forth. The experimental conditions deﬁne the constraint(s) from the conservation laws adapted to the physical situation (conserved total energy, conserved number of particles, etc.). In this chapter, the most usual statistical ensembles will be presented : “microcanonical” ensemble in § 2.2, “canonical” ensemble in § 2.4, “grand canonical” ensemble in § 2.5, using the names introduced at the end of the 19th century. In all these statistical ensembles, one reproduces the considered system in a thought experiment, thus realizing an assembly of macroscopic objects built under the same initial conditions, which allows average values of ﬂuctuating physical parameters to be deﬁned. For a given physical condition, one will decide to use the statistical ensemble most adapted to the analysis of the problem. However, as will be shown, for systems with a macroscopic number of particles (the only situation considered in this book), there is equivalence, to a extremely small relative ﬂuctuation, between the various statistical ensembles that will be described here : in the case of a macroscopic system, a measurement will not distinguish between a situation in which the energy is ﬁxed and another one in which the temperature is given. It is only the convenience of treatment of the problem that will lead us to choose one statistical ensemble rather than another. (On the contrary for systems with a small number of particles, or of size intermediate between the microscopic and macroscopic ranges, the socalled “mesoscopic systems,” nowadays much studied in Condensed Matter Physics, the equivalence between statistical ensembles is not always valid.)

2.1

Determination of the Energy States of an N -Particle System

For an individual particle i in translation, of momentum pi and mass m, under a potential energy Vi (ri ), the time-independent Schroedinger equation (eigenvalues equation) is Ç ˆ i |ψ αi = h i =

å pˆi 2 + Vi (ri ) |ψiαi 2m

(2.1)

αi i εα i |ψi

where αi expresses the diﬀerent states accessible to this particle. As soon as several particles are present, one has to consider their interactions, which leads to a several-body potential energy. Thus, the hamiltonian for N particles in motion, under a potential energy and mutual interactions, is given

Energy States of an N -Particle System

33

by ˆ = H

pˆi 2 i

2m

+ Vi (ri ) +

Vij (ri , rj )

(2.2)

i<j

This is the case, for example, for the molecules of a gas, conﬁned within a limited volume and mutually interacting through Van der Waals interactions ; this is also the situation of electrons in solids in attractive Coulombic interaction with the positively charged nuclei (Vi term), and in repulsion between themselves (Vij term). The latter interactions, expressed in the Vij ’s, are essential in the approach toward thermal equilibrium through the transitions they induce between microscopic states, but they complicate the solution of the N -particle eigenstate problem. Now they mostly correspond to energy terms very small with respect to those associated with the remaining part of the hamiltonian. An analogous problem is faced when one is concerned by the magnetic moments of a ferromagnetic material : the hamiltonian of an individual electronic intrinsic magnetic moment located in a magnetic ﬁeld is given by ˆ = − h µˆB · B

(2.3)

ˆ by : where the Bohr magneton operator µˆB is related to the spin operator S e ˆ µ ˆB = − S m

(2.4)

and has for eigenvalues ∓e/2m (see a course of Quantum Mechanics). The ha contains, miltonian of an ensemble of magnetic moments in a magnetic ﬁeld B ˆ other terms expressing their mutual interin addition to terms similar to h, actions, i.e. ˆ =− + µˆBi · B ˆBi · µˆBj H Jij µ (2.5) i

i<j

where Jij is the coupling term between the moments localized at sites i and j. According to the site i or j, the moment orientation is diﬀerent. All these types of hamiltonians, in which several-particle terms appear, are generally treated through approximations, which allow them to be reduced to ˆ is written as a sum of similar hamiltonians, the simplest situation where H each concerning a single particle. There are several methods to achieve such a reduction : – either neglecting the interactions between particles in the equilibrium state, as they correspond to very small energies. As a consequence, in the above example of the gas molecules, the only term included in the potential energy

34

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

expresses the conﬁnement with a box (conﬁnement potential) or an external ﬁeld (gravity) ; the particles are then considered as “free” ; – or using the so-called “mean ﬁeld” treatment : each particle is subjected to a mean eﬀect from the other ones. This will be a repulsive potential energy in the case of the Coulombic interaction between electrons in a solid, an eﬀective magnetic ﬁeld equivalent to the interactions from the other magnetic moments in the case of a ferromagnetic material ; ˆ in order to decouple the – or changing of variable in the hamiltonian H, ˆ variables, so that H may be re-written as a sum of terms (“normal modes” deﬁning quasi-particles, for example, in the problem of the atomic vibrations in a crystal at a given temperature). From now on, let us assume that the total-system hamiltonian is indeed expressed as the sum of similar individual-particle hamiltonians ˆ i ψ αi (ri ) = εαi ψ αi (ri ) ˆ i , with h ˆ = (2.6) h H i i i i

Then it will only be necessary to solve the problem for an individual particle : indeed you learnt in Quantum Mechanics that the solution of ˆ r1 , r2 , . . . , rN ) = Eψ(r1 , r2 , . . . , rN ) Hψ(

(2.7)

is

ˆj h

N

j

i=1

N α ψiαi (ri ) = εj j ψiαi (ri ) j

(2.8)

i=1

i.e., E=

i

i εα r1 , r2 , . . . , rN ) = i , ψ(

N

ψiαi (ri )

(2.9)

i=1

The total energy of the system is the sum of the individual particle energies and the N -particle eigenfunction is the product of the eigenfunctions for each single particle. In the ﬁrst three chapters of this book, the problems studied will concern discrete variables (for example„ spin magnetic moments) or particles distinguishable by their ﬁxed coordinates, like the atoms of a solid vibrating around their equilibrium position at ﬁnite temperature. The systems with translation degrees of freedom will be analyzed in the classical framework in chapter 4 (ideal gas), or in the framework of the Quantum Statistics from chapter 5 until the end of the course : in the Quantum Statistics description of the accessible

Isolated System in Equilibrium : “Microcanonical Ensemble”

35

microstates, one has to account for the indistinguishability of the particles, as expressed by the Pauli principle. In all that follows, we will assume that the microscopic states of the considered system are known. We are going to rely on the hypothesis of maximum entropy, as introduced in chapter 1, § 1.3 and § 1.4, under the speciﬁc physical constraints due to the given experimental conditions ; we will then deduce the system macroscopic properties, using Statistical Physics.

2.2

Isolated System in Equilibrium : “Microcanonical Ensemble”

In such a system one assumes that there is no possible exchange with its surroundings (Fig. 2.1) : the volume is ﬁxed, the number of particles N is given, the total energy assumes a ﬁxed value E, i.e., its lies in the range between E E + δE, where δE is the uncertainty. This situation is called “microcanonical ensemble.” It is the framework of application of the Boltzmann hypothesis of equiprobability of the W (E) accessible microstates (§ 1.3.3). The probability to realize any of these microstates is equal to 1/W (E) ; the statistical entropy S is maximum under these conditions (with respect to the situation where the probabilities would not be equal) and is given by S = kB ln W (E)

(2.10)

E to δE

Fig. 2.1 : Isolated system. Its energy has a ﬁxed value. To solve this type of problem, one has to determine W (E) : this means performing a combinatory calculation of the number of occurrences of the macroscopic energy E from the various conﬁgurations of the microscopic components of the system. For example, for an ensemble of localized magnetic moments located in an external magnetic ﬁeld, to each ﬁxed value of E corresponds a value of the macroscopic magnetization. The corresponding number of microstates W (E) was evaluated in § 1.3.1. The other physical macroscopic parameters

36

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

are deduced from S and its partial derivatives (see § 3.5). In particular a microcanonical temperature is deﬁned from the partial derivative of S versus E : 1/T = ∂S/∂E.

2.3

Equilibrium Conditions for Two Systems in Contact, Only Exchanging Energy

E A

A ∂S = kB β ∂E

∂S = kB β ∂E

Fig. 2.2: The systems A and A only exchange energy, the combined system A0 =A+A is isolated. Two distinct systems A and A are brought into thermal contact. The system A0 , made up of the combination of A and A (Fig. 2.2) is the isolated system, to which we can apply the above approach, i.e. look for the maximum of its statistical entropy. The systems in contact, A and A , can only exchange energy and are weakly coupled, which means that a possible interaction energy is neglected. We can also say that the system A under study has the constraint to be coupled to A , the total energy E + E being conserved. This is a very common situation, like the one of a drink can in a refrigerator, the ensemble (can-refrigerator) being taken as thermally insulated from the room in which the refrigerator stands, the room representing the surrondings.

2.3.1

Equilibrium Condition : Equal β Parameters

The total energy and the total statistical entropy are shared between A and A : E0 = E + E SA0 (E0 , E) = SA (E) + SA (E0 − E)

(2.11) (2.12)

From the Boltzmann relation (2.10), the equation (2.12) is equivalent to W0 (E0 , E) = WA (E) · WA (E0 − E)

(2.13)

Equilibrium Conditions for Two Systems in Contact

37

since for weakly coupled systems the number of microstates of the combined system is the product of the numbers of microstates of the two subsystems. After a long enough time has elapsed, the combined system (can-refrigerator) A0 reaches an equilibrium situation. Since the combination A0 of the two systems A and A is an isolated system, in equilibrium its entropy is maximum with respect to the exchange of energy between A and A : SA0 (E0 , E) = SA (E) + SA (E0 − E)

(2.14)

∂SA0 (E0 , E) ∂SA0 (E0 , E) ∂SA (E) ∂SA (E0 − E) = 0 , i.e., = + =0 ∂E ∂E ∂E ∂E (2.15) Once the equilibrium has been reached : ∂SA (E0 − E) ∂SA (E ) ∂SA (E) =− = (2.16) ∂E ∂E ∂E The obtained equilibrium condition is that the partial derivative of entropy versus energy takes the same value in both systems in contact, A and A , all the other parameters (volume, etc.) being kept constant. This derivative will be written : ∂S = kB β (2.17) ∂E where kB is the Boltzmann constant. ˜ one In equilibrium, reached for an energy value of system A equal to E = E, thus has β = β

(2.18)

We know that the entropy S and the energy E are both extensive* parameters : by deﬁnition the values of extensive parameters are doubled when the system volume Ω is doubled together with the number N of particles, while the density N/Ω, an intensive* parameter, is maintained in such a transformation. In such a process, the parameter β is not changed, it is an intensive parameter having the dimension of reciprocal energy, the dimension of kB β being a reciprocal temperature. In § 3.3 the partial derivative of the entropy versus energy will be identiﬁed with 1/T , where T is the absolute temperature, so that (2.18) expresses that in equilibrium T = T .

2.3.2

Fluctuations of the Energy Around its Most Likely Value

The number of microstates of the total system A0 realizing the splitting of the energies [E in A, E0 − E in A ] is proportional to the product

38

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

WA (E) · WA (E0 − E). The probability of occurrence of the corresponding microstate is WA (E)WA (E0 − E) WA (E)WA (E0 − E) = p(E, E0 − E) = Wtot (E0 ) WA (E)WA (E0 − E)

(2.19)

E

Let us consider the variation of this probability versus E around the value ˜ achieving the entropy maximum with respect to an energy exchange betE ween A and A . In (2.19), only the numerator is a function of E, since the denominator is a sum over all the possible values of E. For a macroscopic system, W (E) increases very fast with E, as already veriﬁed on the example of free particles in Appendix 3 of Chapter 1. On the other hand, WA (E0 − E) strongly decreases with E if A is macroscopic. Their product goes through a very steep maximum, corresponding to the ˜ state of maximum probability of the combined system, reached for E = E. This energy is very close to the average value E of the energy in system A, ˜ as the probability practically vanishes outside the immediate vicinity of E. ˜ The probability variation around E is much faster than that of its logarithm, which varies versus E in the same way as SA0 (E) = kB ln[WA (E)WA (E0 − E)]

(2.20)

because the combined system A0 is isolated. Note that in the particular case of Å a free-particle system, where W (E) ∼ E N ã ∂S (see Appendix 3, Chapter 1), β = is of the order of N/E. Since the ∂E Ω ˜ the average energy per particle is of the order total energy is of the order E, 1/β = kB T , to a numerical factor of order unity. The average energy of a free particle at temperature T is thus of order kB T . Now we look for the probability variation versus the energy of system A ˜ The logarithm of this probability is for E close to the most likely value E. proportional to the entropy of the combined system SA0 (E), considered as a function of E, as A0 is an isolated system to which the Boltzmann formula SA0 (E) = kB ln W0 (E) applies. In equilibrium, the entropy of the combined system A0 is maximum with respect to an energy exchange between the two systems ; the entropy of A0 ˜ small : can be developed around this maximum, for E − E ã Å 2 Å ã ˜ + ∂SA0 ˜ + 1 ∂ S A0 ˜ 2 (2.21) (E − E) (E − E) SA0 (E) = SA0 (E) ∂E E˜ 2 ∂E 2 E˜

˜ 3 + O (E − E)

Equilibrium Conditions for Two Systems in Contact

39

Here the ﬁrst derivative is equal to zero and the second derivative is negative, as the extremum is a maximum. This second derivative includes the contributions of both A and A and is equal to : ∂ 2 SA ∂ 2 S A ∂ 2 S A0 = + = kB 2 2 ∂E ∂E ∂E 2

Å

∂β ∂β + ∂E ∂E

ã = −kB λ0

(2.22)

˜ In fact, since β = 1/kB T , staAll these terms are calculated for E = E. ting that ∂β/∂E is negative means that the internal energy increases with temperature : this is always realized in practical cases.1 . From the entropy properties (2.21) and (2.22), using the Boltzmann relation one deduces those of the number of accessible microstates : ˜ exp − λ0 (E − E) ˜ 2 /2 W0 (E) ≈ W0 (E)

(2.23)

The probability of realizing the state with the sharing of the energies [E in system A and E0 − E in system A ] is proportional to W0 (E). It is given by

˜ E0 − E) ˜ exp − λ0 (E − E) ˜ 2 /2 p(E, E0 − E) ≈ p(E,

(2.24)

˜ This √ is a Gaussian function around the equilibrium state E. Its width ∆E = 1/ λ0 depends of the size of the total system like Å

∂β ∂β + ∂E ∂E

ã−1/2 (2.25)

that is, like the square root of a number of particles, since E and E are ˜ extensive and the parameters β and β intensive. The relative width ∆E/E varies as a reciprocal number of particles. For a macroscopic system, where N and N ∼ 1023 , this relative width is of the order of 10−11 , thus unmeasurable : the ﬂuctuations of the energy of system A around its equilibrium value are practically null. This is equivalent to saying that, for a system with a very large number of particles, situations i), in which the energy of A is exactly ˜ and ii), in which system A reached an equilibrium equal to the value E, ˜ energy E through exchanges with a heat reservoir, lead to the same physical measurement of the energy. 1 When the total energy has an upper bound, as in the case of the spins in a paramagnetic solid, the maximum energy is reached when all the magnetic moments are antiparallel to the applied magnetic ﬁeld. In the vicinity of this maximum, adding energy produces a decrease in entropy, so that (2.17) leads to a negative β, thus to a “ negative temperature.” However in the situation of a real paramagnetic solid, the spin degree of freedom is in equilibrium with the other degrees of freedom, like the ones related to the atoms’ vibrations around their equilibrium positions. For the latter ones, the energy is not limited and for the whole solid, the total energy indeed increases with temperature.

40

2.4

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

System in Contact with a Heat Reservoir, “Canonical Ensemble”

When two systems are brought into thermal contact, one being much larger than the other (the can is much smaller than the refrigerator) (Fig. 2.3), the larger system behaves like a heat reservoir of energy E0 − E close to E0 , the energy E of the other system thus being very small. The statistical entropy, and, consequently, the parameter β of the larger system, are close to those of the total system : Å Å 2 ã ã ã Å ∂SA (E) ∂ SA (E) ∂SA = − E + O(E 2 ) (2.26) ∂E E0 −E ∂E ∂E 2 E0 E0 Å 2 ã ∂ SA (E) + O(E 2 ) (2.27) = kB β0 − E ∂E 2 E0

E A

A β

Fig. 2.3: The system A is in contact with the heat reservoir A which dictates its parameter β to A.

When the equilibrium is reached, the parameter β of the smaller system is adjusted to the parameter β of the larger system, which itself takes a value close to that of the parameter β0 of the ensemble : the larger system acts as a heat reservoir or thermostat, a body that keeps its value of β and dictates this parameter to the smaller system (we will see in chapter 3 that this is equivalent to saying that it keeps its temperature when put in thermal contact : for example, in equilibrium the can takes the temperature of the refrigerator). The smaller system A can be a macroscopic system, suﬃciently smaller than A , or even a single microscopic particle, like an individual localized magnetic moment inside a paramagnetic solid, with respect to which the rest of the macroscopic solid plays the role of a reservoir. The situation in which the system under study is in thermal contact with a heat reservoir that gives its temperature T to the system is called “canonical,” since it is a very frequent situation among the studied problems. The ensemble

Canonical Ensemble

41

of systems in contact with the heat reservoir, on which we will calculate the ensemble average, is the “canonical ensemble.”

2.4.1

The Boltzmann Factor

We are thus in the situation where a smaller system A is coupled to a larger system A that dictates its value of the parameter β, which is close to the value for the combined isolated system. First consider the probability pi that a speciﬁc microstate i of system A, of energy Ei , is produced. It only depends on the properties of the heat reservoir and is the ratio of the number of microstates Wres (E0 − Ei ) of the reservoir with this energy to the sum of all the numbers of microstates for all the reservoir energies E0 − Ei : Wres (E0 − Ei ) pi = Wres (E0 − Ei )

(2.28)

i

Wres (E0 −Ei ) is calculated in the microcanonical ensemble for the given energy E0 − Ei and is related to Sres (E0 − Ei ) through the Boltzmann relation ln Wres (E0 − Ei ) =

1 Sres (E0 − Ei ) kB

(2.29)

Indeed, once the reservoir energy E0 − Ei is ﬁxed, the diﬀerent microstates in this situation are equally likely. Since

Å Sres (E0 − Ei ) = SA (E0 ) − Ei

∂SA ∂E

ã + ...

(2.30)

E =E0

ln Wres (E0 − Ei ) = ln WA (E0 ) − Ei β0 + . . .

(2.31)

= constant − β0 Ei + . . . the probability of occurrence of this microstate is proportional to Wres (E0 − Ei ), i.e., pi = C exp(−β0 Ei )

(2.32)

In the probability pi there appears the exponential of the energy Ei of the system. This is the so-called “Boltzmann factor” (1871), or “canonical distribution,” an expression that will be used very often in this course, every time we will study an ensemble of distinguishable particles, in ﬁxed number 1 (the case of indistinguishable particles will N , of given temperature T = kB β

42

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

be treated in the framework of the Quantum Statistics in chapter 5 and the following ones). As an example, Fig. 2.4 schematizes the Boltzmann factor for a 4-microstate system. E

E β1 E4

E3

E2

E1

β2 = 0, 5β1 (T2 = 2T1 )

p4

E4

p3

p4

E3

p2

E2

p1

E1

p3

p2

p1

Fig. 2.4: The considered system has four microstates of energies E1 to E4 . The abscissae are proportional to the probabilities of occurrence of one of the microstates, for two diﬀerent values of β (or of the temperature). If the total number of microstates of system A corresponding to the same energy E is W (E), then the probability that system A, coupled to A , has the energy E, is equal to : p(E) = C W (E) exp(−β0 E)

(2.33)

where W (E) is the degeneracy (the number of ways to realize this energy) of the state of energy E. The probability p(E) is the product of the function W (E), very rapidly increasing with energy for a macroscopic system, by a ˜ decreasing exponential : the probability has a very steep maximum for E = E.

2.4.2

Energy, with Fixed Average Value

We now show that another condition, i.e., the constraint of a given average value E of the energy of a system A, leads to a probability law analogous to

Canonical Ensemble

43

the one just found. The only diﬀerence is that the value of β, now a parameter, should be adjusted so that the average of energy indeed coincides with the given value E. We have to maximize the statistical entropy of A while the value E is given. We thus have to realize SA = −kB

pi ln pi maximum, under the constraints

i

⎧ ⎪ pi = 1 ⎪ ⎨ i ⎪ pi Ei = E ⎪ ⎩

(2.34)

i

A general mathematical method allows one to solve this type of problem, it is the method of Lagrange multipliers explained in Appendix 2.1 : parameters are introduced, which are the Lagrange multipliers (here β). These parameters are adjusted at the end of the calculation, in order that the average values be equal to those given by the physical properties of the system, here the value of the average energy. One thus obtains for the probability of occurrence of the microstate of energy Ei : exp(−βEi ) pi = exp(−βEi )

(2.35)

i

that is, an expression of the “Boltzmann-factor” type (2.32) but here β does not characterize the temperature of a real heat reservoir. It is rather determined by the condition that the average energy of the system A should be the one given by the problem conditions, that is :

Ei exp(−βEi )

i

exp(−βEi )

= E

(2.36)

i

This method of the Lagrange multipliers is followed in some Statistical Physics courses to obtain the general expression of the Boltzmann factor (for example, those given at Ecole polytechnique by R. Balian, E. Brézin, or A. Georges and M. Mézard).

44

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

Partition Function Z

2.4.3

The probability for the system in contact with a heat reservoir to be in a state of energy Ei is pi =

e−βEi 1 , withβ = ZN kB T

The term ZN =

(2.37)

e−βEi

(2.38)

i

which allows one to norm the Boltzmann factors is called the “canonical partition function for N particles” (in German “Zustandssumme,” which means “sum over the states”). Formulae (2.37) and (2.38) are among the most important ones of this course ! The complete statistical information on the considered problem is contained in its partition function, since from Z and its partial derivatives all the average values of the physical macroscopic parameters of the system can be calculated. We will show this property now on the examples of the average energy and the energy ﬂuctuation around its average (for the entropy see § 2.4.5). Indeed, the average energy is given by E =

pi Ei =

i

Ei

i

exp(−βEi ) Z

(2.39)

∂ ln Z 1 ∂Z =− =− Z ∂β ∂β It appears here that ln Z is extensive like E.

To determine the energy ﬂuctuation around its average value, one calculates (∆E)2 = E − E2 = E 2 − E2

(2.40)

A procedure similar to that of (2.39) is followed to ﬁnd E 2 : E 2 =

pi Ei2 =

i

(∆E)2 =

1 ∂2Z 1 − 2 Z ∂β 2 Z

Ei2

i

Å

∂Z ∂β

exp(−βEi ) 1 ∂2Z = Z Z ∂β 2 ã2 =

∂ ∂β

Å

1 ∂Z Z ∂β

ã =

(2.41) ∂ 2 ln Z ∂β 2

(2.42)

One veriﬁes here that, since ln Z varies as the number of particles of the system and β is intensive, ∆E varies as N 1/2 and ∆E/E as N −1/2 (in agreement

Canonical Ensemble

45

with the result of § 2.3.2), which corresponds to an extremely small value in a macroscopic system : for N = 1023 , ∆E/E is of the order of 10−11 ! Note 1 : A demonstration (outside the framework of the present course, see, for example, R. Balian chapter 5 § 5) can be done of the equivalence between all the statistical ensembles in the case of macroscopic systems : it is valid in the “thermodynamical limit,” which consists in simultaneously having several parameters tending toward inﬁnity : the volume Ω (which means in particular that it is very much larger than typical atomic volumes), the particle number N and the other extensive parameters, while keeping constant the particle density N/Ω and the other intensive parameters. In such a limit, it is equivalent to impose an exact value of a physical parameter (like the energy in the microcanonical ensemble) or an average value of the same parameter (the energy in the canonical ensemble). In fact, the technique is simpler in the canonical ensemble, where ln Z is calculated on all the states without any restriction, than in the microcanonical ensemble, where the calculation of ln W (E) requires the limitation to the range of energies between E and E + δE. Besides, for macroscopic systems one understands that the predicted physical results, using either the microcanonical or the canonical statistical ensemble, cannot be distinguished through measurements. Note 2 : The value of ZN “gives a hint” at the number of microstates that can be achieved at the experiment temperature. Indeed at very low temperature, where βEi → ∞, only the fundamental state E0 is realized and ZN = 1. On the other hand, at high temperature where βEi → 0, many terms are of the order of unity, corresponding to practically equally likely states.

2.4.4

Entropy in the Canonical Ensemble

We just showed that, using the partition function, we can calculate the average energy of system A at temperature T , together the ﬂuctuation around this average energy. We now have to express the statistical entropy of A in these conditions. Here the probabilities of occurrence of the diﬀerent microstates are not equal, since they depend of their respective energies through Boltzmann factors. Thus the deﬁnition of the statistical entropy to be used is no longer that of Boltzmann, but rather the one of Gibbs : SA = −kB

i

pi ln pi

(2.43)

46

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

with pi =

1 −βEi e , pi = 1 Z i

(2.44)

Consequently, SA = −kB

pi (− ln Z − βEi )

(2.45)

i

SA = kB (ln Z + βE) = kB ln Z + kB βU

(2.46)

The average energy E of the macroscopic system is identiﬁed to the internal energy U of Thermodynamics (see § 3.2). Note that, for a large system in which the energy ﬂuctuation is relatively very small around E, the probability of occurrence of this energy value is very close to unity. The number of occurrences of this average energy is W (E), so that the Gibbs entropy (2.43) practically reduces to the Boltzmann entropy (2.10) for this value E.

2.4.5

Partition Function of a Set of Two Independent Systems with the Same β Parameter (Same T )

In § 2.1 we saw that, in the N -particle problems solved in this book, we always decompose the hamiltonian of the total system into a sum of hamiltonians for individual particles ; a particle may have several degrees of freedom, whence another sum of hamiltonians. The energy of a microstate is thus a sum of energies and now we see the consequence of this property on the partition function of such a system in thermal equilibrium. This situation is schematized by simply considering a system A made up of two independent subsystems A1 and A2 , both in contact with a heat reservoir at temperature T . A particular microstate of A1 has the energy E1i , while the energy of A2 is E2j and the energy of A is Eij . Then ⎧ Eij = E1i + E2j ⎪ ⎪ ⎪ ⎪ ⎪ ⎪ 1 ⎪ ⎪ Z = Z1 Z2 = exp{−β(E1i + E2j )} , with β = ⎪ ⎪ kB T ⎨ ij ln Z = ln Z1 + ln Z2 ⎪ ⎪ ⎪ ⎪ ⎪ ⎪ E = E1 + E2 ⎪ ⎪ ⎪ ⎪ ⎩ S = S1 + S2

(2.47)

Canonical Ensemble

47

This property is very often used : when the energies of two independent systems at the same temperature sum, the corresponding partition functions multiply. Consequently, in the case of distinguishable independent particles with several degrees of freedom, one will separately calculate the partition functions for the various degrees of freedom of an individual particle ; then one will multiply the factors corresponding to the diﬀerent degrees of freedom and particles to obtain the partition function Z of the total system. Finally, in the special case of the ideal gas, one will introduce the factor CN into Z (see § 4.4.3 and 6.7). Take the example of the vibrations of atoms around their equilibrium positions, owing to thermal motion, in a solid in thermal equilibrium at temperature T : this is a very classical exercise and here we only sketch its solution. In the model proposed by Einstein (1907), one assumes that the atoms, in number N , are points and that they are all attracted toward their respective equilibrium position with the same restoring constant, that is, with the same frequency ω. The value of ω depends of the mechanical properties of the solid, in particular of its stiﬀness. Then the hamiltonian for a particular atom in motion is given by 1 ˆ 2i ˆ xi + ˆhyi + h ˆ zi ˆi = p + mω 2ri2 = h h 2m 2

(2.48)

It is the sum of three hamiltonians of the “harmonic-oscillator” type, identical ˆ xi relative to the for each coordinate. The eigenvalues of the hamiltonian h coordinate x of site i are

1 n = nxi + ω Exi 2

(2.49)

The contribution to the partition function of this degree of freedom is equal to exp(−β(nxi + 1/2)ω) (2.50) zxi = n

The partition function for the total system is the product of 3N terms similar to this one. The average energy is deduced using (2.39), it is equal to 3N times the average value for the coordinate x of site i. The parameter accessible to experiment is the speciﬁc heat at constant volume Cv , deﬁned as the derivative of the average energy with respect to temperature. The above solution gives for this lattice contribution Cv =

3N kB (Θ/T )2 ω , taking =Θ Θ/T 2 kB (e − 1)

This expression is sketched on Fig. 2.5.

(2.51)

48

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

Cv /3N kB 1

O

1

T /Θ

Fig. 2.5: Variation of Cv versus temperature, according to the Einstein model. The dots are experimental data for diamond, the solid line is calculated for Θ = 1320 K (after A. Einstein, Annalen der Physik vol.22, p.186 (1907)). Its high temperature limit Cv = 3N kB = 25 J.K−1mol−1 is independent of ω, thus universal : this is the Dulong and Petit law (1819), valid for metals at room temperature. The low temperature limit Cv → 0 for T → 0 is in agreement with experiment. However, the above law of variation of Cv is not experimentally veriﬁed. The Einstein model is improved by introducing a repartition of characteristic frequencies in a solid instead of a single ω : this is the Debye model (1913), in which Cv is proportional to T 3 at very low temperatures. The agreement with experiment then becomes excellent for insulators, whereas for metals an additional term is due to the mobile electron contribution to the speciﬁc heat (see § (7.2.2)).

2.5

Exchange of Both Energy and Particles : “Grand Canonical Ensemble”

Here we consider two coupled systems A and A , separated by a porous partition through which energy and particles can be exchanged, the combined system A0 being isolated from its surroundings (Fig. 2.6). An example of such a situation is water molecules under two phases, liquid and gas, contained in

Grand Canonical Ensemble

∂S = −kB α ∂N A ∂S = kB β ∂E

49

E

∂S = −kB α ∂N A

N

∂S = kB β ∂E

Fig. 2.6 : The two systems A and A exchange energy and particles. a thermally insulated can of ﬁxed volume, the water molecules belonging to either phase. One can also study hydrogen molecules as a free gas in equilibrium with hydrogen molecules adsorbed on a solid palladium catalyst surface together with the total system being thermally insulated and the total number of molecules ﬁxed. We have to state that the entropy is maximum at equilibrium, following an approach very similar to the one of 2.3.1.

2.5.1

Equilibrium Condition : Equality of Both Temperatures and Chemical Potentials

One writes that the maximum of entropy for the combined system A0 = A+A takes place at constant total energy and ﬁxed total number of particles, i.e., SA0 (E0 , N0 ) = SA (E, N ) + SA (E , N ) maximum under the two constraints

(2.52)

⎧ ⎨ E0 = E + E

(2.53)

⎩ N = N + N 0 Let us express the condition of entropy maximum of A0 : SA0 (E0 , N0 ) = SA (E, N ) + SA (E0 − E, N0 − N )

(2.54)

With respect to an energy exchange between the two systems : Å ã ∂SA0 (E, N ) =0, ∂E N Å i.e.,

∂SA0 (E, N ) ∂E

ã

Å =

N

∂SA (E, N ) ∂E

ã

Å +

N

∂SA (E0 − E, N0 − N ) ∂E

ã =0 N

(2.55)

50

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

With respect to a particles exchange : Å ã ∂SA0 (E, N ) =0 ∂N E Å

∂SA0 (E, N ) i.e., ∂N

ã

Å =

E

∂SA (E, N ) ∂N

ã

Å + E

∂SA (E0 − E, N0 − N ) ∂N

ã = 0. E

(2.56) The thermal equilibrium condition (2.55) again provides the results of § 2.3.1, that is, β = β . The condition (2.56) on N implies that ã ã ã Å Å Å ∂SA (E0 − E, N0 − N ) ∂SA (E , N ) ∂SA (E, N ) =− = ∂N ∂N ∂N E E E (2.57) The second equilibrium condition is the equality, in the systems A and A in contact, of the partial derivatives of the entropy versus the number of particles, all the other parameters being kept constant. This derivative will be written Å ã ∂S = −kB α (2.58) ∂N Ω,E where kB is the Boltzmann constant. It will be seen in § 3.5.1 that the parameter α is related to the chemical potential µ through α = βµ = µ/kB T . The equilibrium condition between systems A and A thus simultaneously implies ⎧ ⎪ ⎨ β = β , i.e., T = T (2.59) ⎪ ⎩ α = α , i.e., µ = µ T T that is, the equality of both the temperatures and the chemical potentials in both systems. Thus in equilibrium the chemical potential of hydrogen is the same, whether in gas phase or adsorbed on the catalyst, and the temperatures are the same in both phases.

2.5.2

Heat Reservoir and Particles Reservoir ; Grand Canonical Probability and Partition Function

Similarly to the approach of § 2.4, one now considers that the system A under study is macroscopic, but much smaller that the system A (Fig. 2.7). Consequently, the parameters α and β of system A are almost the same as those

Grand Canonical Ensemble

51

of the combined system A0 , and are very little modiﬁed when system A varies in energy or in number of particles : with respect to A the system A behaves like a heat reservoir and a particle reservoir, it dictates both its temperature β = 1/kB T and its chemical potential α = µ/kB T to A.

α

E A

A N

β

Fig. 2.7: The system A is in contact with the heat reservoir and particles reservoir A , which dictates its parameters α and β to A. This situation in Statistical Physics, in which the system under study A is coupled to a heat reservoir that dictates its value of β, i.e., its temperature, and to a particle reservoir that gives the value of α, is called the “grand canonical ensemble.”

2.5.3

Grand Canonical Probability Law and Partition Function

Now both the energy Ei and the number of particles Ni of the considered system enter into the Boltzmann-type probability pi . Indeed, consider the probability pi to reach a particular microstate i of A, of energy Ei and particle number Ni . It is related to the ability for the reservoir to realize this physical situation at given energy and number of particles, that is, to the number Wres (E0 − Ei , N0 − Ni ) of equally likely microstates, and thus to the reservoir entropy, through the Boltzmann entropy relation. Now Å ã ∂Sres Sres (E0 − Ei , N0 − Ni ) = SA0 (E0 , N0 ) − Ei ∂E E0 ,N0 Å ã (2.60) ∂Sres − Ni + ... ∂E E0 ,N0 thus ln Wres (E0 − Ei , N0 − Ni ) = ln Wres (E0 , N0 ) − Ei β0 + Ni α0 + . . .

(2.61)

The probability of occurrence of this microstate is proportional to Wres , i.e., p(Ei , Ni ) =

1 exp(−β0 Ei + α0 Ni ) ZG

(2.62)

52

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

which deﬁnes the grand partition function ZG , the quantity which norms the probabilities. It is now a sum over all energies and numbers of particles that can be achieved ; this sum can also be regrouped according to the diﬀerent numbers of particles : e−βEi +αNi = eαN ZN (2.63) ZG = i

N

The canonical partition function ZN for N particles, as deﬁned in (2.38), e−βEi (2.64) ZN = i

thus appears in ZG with a weight equal to exp αN . (Note that in (2.63) the sum is performed over all microstates i with an arbitrary number of particles, whereas in (2.64) the sum is only over the states with N particles, with N ﬁxed.)

2.5.4

Average Values

Using p(Ei , Ni ) one calculates the average energy and the average number of particles of system A : this is conveniently performed from the partial derivatives of ZG . The approach is the same as in (2.39) : E =

Ei p(Ei , Ni ) =

i

Ei

i

∂ ln ZG exp(−βEi + αNi ) 1 ∂ZG =− =− ZG ZG ∂β ∂β (2.65)

N =

Ni p(Ei , Ni ) =

i

i

Ni

∂ ln ZG exp(−βEi + αNi ) 1 ∂ZG = =+ ZG ZG ∂α ∂α (2.66)

To determine the energy ﬂuctuations around its average value, one proceeds in a way similar to the calculation in the canonical ensemble (eqs. (2.41) and (2.42)). Let us calculate the ﬂuctuation of the particle number in system A : N 2 =

i

Ni2 p(Ei , Ni ) =

1 ∂ 2 ZG ZG ∂α2

1 ∂ 2 ZG − (∆N )2 = N 2 − N 2 = ZG ∂α2 Å ã 1 ∂ZG ∂ 2 ln ZG ∂ = = ∂α ZG ∂α ∂α2

Å

(2.67) 1 ∂ZG ZG ∂α

ã2 (2.68)

Other Statistical Ensembles

53

Since ln ZG is extensive as it increases like E or N (see (2.65) and (2.66)), (∆N )2 varies proportionaly with the size of the system, or with its average number of particles. Consequently, ∆N/N is of the order N −1/2 , it is thus very small for a macroscopic system. Thus by diﬀerentiating ZG one obtains the average physical parameters E and N and their ﬂuctuations around these averages : in fact ZG , which expresses the probabilities, contains the complete information on the system.

2.5.5

Grand Canonical Entropy

Using the Gibbs deﬁnition S = −kB

pi ln pi

(2.69)

i

one calculates the entropy in the grand canonical ensemble. Here pi =

1 −βEi +αNi e , pi = 1 ZG i

(2.70)

Consequently, S = −kB

pi (− ln ZG − βEi + αNi )

(2.71)

i

S = kB (ln ZG + βE − αN ) = kB ln ZG + kB βU − kB αN

(2.72)

In the present deﬁnition of the grand canonical ensemble, we have considered a macroscopic system in contact with a heat reservoir and a particle reservoir. The grand canonical ensemble is also used when the system under study is speciﬁed by the average values of both its energy and its particle number and one then needs to determine β and α = βµ : the argument is the extension of the one in § 2.4.2. We will use the grand canonical ensemble in particular for the study of the Quantum Statistics : the Pauli principle introduces limitations to the occupation of the quantum states, which will be more conveniently expressed in this ensemble.

2.6

Other Statistical Ensembles

As you imagine, one can generalize to various physical situations the approach consisting of considering two weakly coupled systems A and A , the combined system being isolated, i.e., the total parameters being ﬁxed. In particular one

54

Chapter 2. The Diﬀerent Statistical Ensembles. General Methods

can study systems A and A exchanging both energy and volume, the total volume Ω0 = Ω + Ω being ﬁxed. The equilibrium condition now implies that the partial derivatives of the entropy ∂S ∂E

N,Ω

= kB β

and

∂S ∂Ω

N,E

= kB γ

(2.73)

take the same value in both systems. Thermodynamics will show, in § 3.5.1, that kB γ = P/T where P is the pressure, so that the new equilibrium condition will be ® T = T (2.74) P = P that is, in equilibrium the systems A and A will have the same temperature and the same pressure. If A is much larger than A, it will give its pressure to A. The chemistry experiments performed in isothermal and isobar conditions (for example, at room temperature and under the atmospheric pressure) are in such a condition. Then Å ã P Ωi 1 1 Ei − exp(−βEi − γΩi ) = exp − (2.75) p(Ei , Ωi ) = ZT,P ZT,P kB T kB T i represents the microstate in which the system has the energy Ei and occupies the volume Ωi . A new partition function ZT,P is deﬁned by exp(−βEi − γΩi ) (2.76) ZT,P = i

(In fact the sum over the volume, a continuous variable, is rather an integral.) The entropy is deduced : S = −kB p(Ei , Ωi )(− ln ZT,P − βEi − γΩi ) (2.77) i

S = kB (ln ZT,P + βE + γΩ)

(2.78)

Summary of Chapter 2 In this chapter the general method for solving problems in Statistical Physics is presented. – First one needs to determine the quantum states of the system. In the cases we will consider, the system hamiltonian can be decomposed into a sum of similar one-particle terms ˆi ˆ = H h i

Until chapter 4, the particles will be assumed to be distinguishable. – Only in a second step does one choose a statistical ensemble adapted to the physical conditions of the problem (or the most convenient to solve the problem, since for macroscopic systems one cannot measure the ﬂuctuations resulting from the choice of a speciﬁc ensemble rather than another one). One expresses that the entropy is maximum in equilibrium, under the constraints set by the physical conditions. According to the chosen ensemble, a mathematical quantity deduced from the probabilities allows one to obtain the macroscopic physical parameters that can be measured in experiments : – the microcanonical ensemble is used to treat the case of an isolated system : the number of microstates W (E) all producing the same macrostate of energy in the range between E and E + δE, allows the entropy to be obtained S = kB ln W (E) – the canonical ensemble is adapted to the case of a system with a ﬁxed number of particles N , in contact with a heat reservoir at temperature T . The probability to be in a microstate of energy Ei is then given by the Boltzmann factor : e−βEi 1 , with β = pi = ZN kB T 55

56

Summary of Chapter 2

The canonical partition function for N particles is deﬁned by ZN = e−βEi i

It is related to the average energy through E = −

∂ ln Z 1 ∂Z =− Z ∂β ∂β

– the grand canonical ensemble corresponds to the situation of a system in contact with a heat reservoir and a particle reservoir. The grand canonical partition function is introduced : ZG = eαN ZN N

The probability to be in a microstate of both energy Ei and number of particles Ni is given by p(Ei , Ni ) =

1 exp(−βEi + αNi ) ZG

The average energy and the average number of particles are then deduced : E = −

∂ ln ZG 1 ∂ZG ∂ ln ZG 1 ∂ZG =− , N = = ZG ∂β ∂β ZG ∂α ∂α

Appendix 2.1 Lagrange Multipliers One looks for the extremum of a function with several variables, satisfying constraints. Here we just treat the example of the entropy maximum with a ﬁxed average value of the energy, i.e., pi ln pi maximum, under the constraints S = −kB i

⎧ pi = 1 ⎪ ⎨ i ⎪ pi Ei = E ⎩

(2.79)

i

The process consists in introducing constants, the Lagrange multipliers, here λ and β, the values of which will be determined at the end of the calculation. Therefore one is looking for the maximum of the auxiliary function : F (p1 , . . . , pi , . . . ) = − pi ln pi + λ( pi − 1) − β( Ei pi − E) (2.80) i

i

i

which expresses the above conditions, with respect to the variables pi . This is written ∂F = −(ln pi + 1) + λ − βEi = 0 ∂pi

(2.81)

ln pi = −βEi + λ − 1

(2.82)

pi = (eλ−1 ) e−βEi

(2.83)

The solution is

Consequently,

57

58

Appendix 2.1

The constants λ and β are then obtained by expressing the constraints : • pi = 1 , whence i

eλ−1

e−βEi = 1

i

eλ−1 = •

1 e

−βEi

=

(2.84) 1 Z

(2.85)

i

pi Ei = E , i.e.,

(2.86)

i

1 Ei e−βEi = E Z i

(2.87)

which is indeed expression (2.36). Note that, if the only constraint is Σpi = 1, the choice of the pi maximizing S i

indeed consists in taking all the pi equal ; this is what is done for an isolated system using the Boltzmann formula.

Chapter 3

Thermodynamics and Statistical Physics in Equilibrium In Chapter 1 we stated the basic principles of Statistical Physics, which rely on a detailed microscopic description and on the hypothesis, for an isolated system, of the equiprobability of the diﬀerent microscopic occurrences of a given macroscopic state ; the concept of statistical entropy was also introduced. In Chapter 2 the study was extended to other systems in equilibrium, in which the statistical entropy is maximized, consistently with the constraints deﬁned by the physical conditions of the problem. It is now necessary to link this statistical entropy S and the entropy which appears in the second law of Thermodynamics and that you became familiar with in your previous studies. It will be shown in this chapter that these two entropies are identical. Thus the purpose of this chapter will ﬁrst be (§ 3.1 to 3.4), mostly in the equilibrium situation, to rediscover the laws of Thermodynamics, which rule the behavior of macroscopic physical properties. This will be done in the framework of the hypotheses and results of Statistical Physics, as obtained in the ﬁrst two chapters of this book. Then we will be able to use at best the diﬀerent thermodynamical potentials (§ 3.5), each of them being directly related to a partition function, that is, to a microscopic description. They are convenient tools to solve the Statistical Physics problems. We are going to state the successive laws of Thermodynamics and to see how each one is interpreted in the framework of Statistical Physics. The example chosen in Thermodynamics will be the ideal gas, the properties of which you 59

60

Chapter 3. Thermodynamics and Statistical Physics

studied in previous courses. Its study in Statistical Physics will be presented in the next chapter 4.

3.1

Zeroth Law of Thermodynamics

According to this law, if two bodies at the same temperature T are brought into thermal contact, i.e. can exchange energy, they remain in equilibrium. In Statistical Physics, we saw in (2.3.1) that if two systems are free to exchange energy, the combined system being isolated, the equilibrium condition is that the partial derivative of the statistical entropy versus energy should take the same value in each system : ∂S ∂S = kB β, = kB β ; β = β (3.1) ∂E ∂E It was suggested that this partial derivative of the entropy is related to its temperature, as will be more precisely shown below in § 3.3.

3.2

First Law of Thermodynamics

Its statement is as follows : Consider an inﬁnitesimal process of a system with a ﬁxed number of particles, exchanging an inﬁnitesimal work δW and an inﬁnitesimal amount of heat δQ with its surroundings. Although both δW and δQ depend on the sequence of states followed by the system, the sum dU = δW + δQ

(3.2)

only depends on the initial state and the ﬁnal state, i.e., it is independent of the process followed. This states that dU is the diﬀerential of a state function U , which is the internal energy of the system. A property related to this ﬁrst law is that the internal energy of an isolated system remains constant, as no exchange of work or heat is possible : dU = 0

(3.3)

In this Statistical Physics course, we introduced the energy of a macroscopic system : in the microcanonical ensemble it is given exactly ; in the canonical or grand canonical ensemble its value is given by the average pi Ei (3.4) E = U = i

First Law of Thermodynamics

61

where pi is the probability of occurrence of the state of energy Ei . The diﬀerential of the internal energy pi dEi + Ei dpi (3.5) dU = i

i

consists in two terms that we are going to identify with the two contributions in dU , i.e., δW and δQ. For this purpose, two particular processes will ﬁrst be analyzed, in which either work (§ 3.2.1) or heat (§ 3.2.2) is exchanged with the surroundings. Then the case of a quasi-static general process will be considered (§ 3.2.3).

3.2.1

Work E

E

E4 + dE4 E4 E3 + dE3 E3

E2 E2 + dE2

E1 + dE1 E1 xj

xj + dxj

Fig. 3.1: Modiﬁcation of the energies of the considered system when the external parameter varies from xj to xj + dxj . The quantum states of a system are the solutions of a hamiltonian which depends of external parameters xj characteristic of the system : for example, you learnt in Quantum Mechanics, and will see again in chapter 6, that a gas molecule, free and conﬁned within a ﬁxed volume Ω, assumes for energy

62

Chapter 3. Thermodynamics and Statistical Physics

eigenvalues 2 k 2 /2m, where the allowed values of k depend of the container volume like Ω−1/3 ; as for the probabilities of realizing these eigenstates, they are determined by the postulates of Statistical Physics. Mechanical work will inﬂuence these external parameters : for example, the motion of a piston modiﬁes the volume of the system, thus shifting its quantized states. More generally, if only the external parameters of the hamiltonian are modiﬁed from xj to xj + dxj , the energy of the eigenstates is changed and a particular eigenvalue Ei of the hamiltonian becomes Ei + dEi in such a process (Fig. 3.1). The variation of the average energy of the system is then ∂U dxj = −Xj dxj ∂xj

(3.6)

where Xj is the generalized force corresponding to the external parameter xj . In the special case of a gas, the variation of internal energy in a dilatation dL is equal to ∂U dL = −F dL = −P SdL = −P dΩ ∂L

(3.7)

Here P is the external pressure that acts on the system surface S ; it is equal to the pressure of the system in the case of a reversible process (Fig. 3.2). This internal energy variation is identiﬁed with the inﬁnitesimal work received by the gas. In such a situation the variation dL of the characteristic length modiﬁes the accessible energies. P dΩ

S

dL

Ω

Fig. 3.2 : Dilatation of the volume Ω submitted to the external pressure P . In general the variation of average energy resulting from the variation of the external parameters is identiﬁed with the inﬁnitesimal work δW = − Xj dxj (3.8) j

and corresponds to the term

i

pi dEi

First Law of Thermodynamics

3.2.2

63

Heat

Assume that energy is brought into the system, while no work is delivered : the external parameters of the hamiltonian are unchanged, the energy levels Ei remain identical. The average energy of the system increases, now due to higher probabilities of occurrence of the system states of higher energy. The probabilities pi are modiﬁed and now the variation of internal energy is identiﬁed with the inﬁnitesimal heat absorbed : δQ = Ei dpi (3.9) i

A way to conﬁrm this identiﬁcation of

Ei dpi is to consider a very slow pro-

i

cess in which no heat is exchanged with the surroundings, that is, the system is thermally insulated : this is an adiabatic ∗ process. In Thermodynamics, such a process is done at constant entropy. In Statistical Physics, it can also be shown that the statistical entropy is conserved in this process : pi ln pi = constant (3.10) S = −kB i

Consequently,

Now as

dpi ln pi +

i

pi = 1, one has

i

pi

i

dpi =0 pi

(3.11)

dpi = 0, whence

i

ln pi dpi = 0

(3.12)

i

One replaces ln pi by its expression in the canonical ensemble ln pi = −βEi − ln Z and deduces

Ei dpi = 0

(3.13)

(3.14)

i

which is indeed consistent with the fact that the heat δQ exchanged with the surroundings is zero in an adiabatic process.

3.2.3

Quasi-Static General Process

Consider an inﬁnitesimal quasi-static ∗ process of a system, that is, slow enough to consist in a continuous and reversible ∗ sequence of equilibrium states.

64

Chapter 3. Thermodynamics and Statistical Physics

In the diﬀerential dU one still identiﬁes the work with the term arising from the variation of the hamiltonian external parameters in this process. The complementary contribution to dU is identiﬁed with δQ, the heat exchanged with the surroundings in the process.

3.3

Second Law of Thermodynamics

Let us recall its statement : Let us consider a system with a ﬁxed number N of particles, in thermal contact with a heat source at temperature T . The system exchanges an amount of heat δQ with this source in an inﬁnitesimal quasi-static process. Then the entropy variation dSthermo of the considered system is equal to dSthermo =

δQ T

(3.15)

Although δQ depends on the particular path followed, the extensive quantity dSthermo is the diﬀerential of a state function, which thus only depends on the initial and ﬁnal states of the considered process. Besides, in a spontaneous, i.e. without any contribution from the surroundings, and irreversible process, the entropy variation dSthermo is positive or zero. The general deﬁnition of the entropy in Statistical Physics is that of Gibbs : S = −kB

pi ln pi

(3.16)

i

Let us ﬁrst examine the interpretation of an irreversible process in Statistical Physics. An example is the free expansion of a gas into vacuum, the so-called “Joule-Gay-Lussac expansion” : a N -molecule gas was occupying a volume Ω (Fig. 3.3a), the partition is removed, then the gas occupies the previously empty volume Ω (Fig. 3.3b). The combined system is isolated and removing the partition requires a negligible amount of work. The accessible volume becomes Ω+Ω and it is extremely unlikely that the molecules again gather in the sole volume Ω. In such a process, which is irreversible in Thermodynamics, the accessible volume for the particles has increased, that is, the constraint on their positions has been removed ; the number of cells of the one-particle phase space (r, p) has increased with the accessible volume in this process. Let us now consider an inﬁnitesimal reversible process, corresponding to a mere modiﬁcation of the probabilities of occurrence of the same accessible

Second Law of Thermodynamics

65

Ω

Ω

Ω

Ω

a)

b)

Fig. 3.3 : Joule-Gay-Lussac expansion into vacuum. microstates. dS is obtained by diﬀerentiating (3.16) : pi ln pi dpi + dpi dS = −kB pi i i ln pi dpi = −kB i

since, as probabilities are normalized,

(3.17)

dpi = 0.

i

To be more precise, let us work in the canonical ensemble, in which the system is submitted to this process at constant number of particles, remaining at every time t in thermal equilibrium with a heat reservoir at the temperature T (t). The probability to realize the microstate of energy Ei is equal to (see (2.37)) : pi (t) =

e−β(t)Ei (t) , i.e., ln pi (t) = −β(t)Ei (t) − ln Z(t) Z

Then dS = kB β(t)

Ei (t)dpi

(3.18)

(3.19)

i

From the above interpretation of the inﬁnitesimal heat δQ, (3.19) can also be written dS = kB β(t)δQ

(3.20)

This expression is to be compared to (3.15). The statistical entropy deﬁned by (3.16) and the thermodynamical entropy (3.15) can be identiﬁed under the condition that kB β, deﬁned by (3.1) as the partial derivative of the statistical 1 entropy S versus energy, is identical to , the reciprocal of the thermodynaT mical temperature : kB β ≡

1 T

(3.21)

66

Chapter 3. Thermodynamics and Statistical Physics

Indeed the integrand factor, which allows one to transform δQ into the total diﬀerential dS, is unique to a constant, which justiﬁes this identiﬁcation.

3.4

Third Law of Thermodynamics

The second law allows one to calculate entropy variations. From the third law, or Walther Nernst (1864-1941)’s law, the absolute value of S can be obtained : When the absolute temperature T of a system tends to zero, its entropy tends to zero. From a microscopic point of view, when the temperature vanishes, the particles gather into the accessible states with the lowest energies. If the minimum energy corresponds to a single state E0 , the probability of occupation of this fundamental state tends to unity while those for the other states vanish. Then the microscopic description becomes certain and the entropy vanishes : for example, when the temperature tends to zero in a monoatomic ideal gas, one expects the molecules to be in a state of kinetic energy as small as possible (yet diﬀerent from zero, because of the Heisenberg uncertainty principle) ; in the case of electrons in a solid, as will be studied in chapter 7, the fundamental state is nondegenerate too. Let us now consider a material which has two equivalent states at low temperature, with the same minimal energy, for example, two allotropic crystalline species, each occurring with the probability 1/2. When T tends to zero, the corresponding entropy tends to S0 = kB ln 2

(3.22)

This quantity is diﬀerent from zero, yet it is not extensive so that S0 /N becomes negligible. This is very diﬀerent from the situation at ﬁnite T where S/N is ﬁnite and S extensive. Thus the statement of the Nernst law consistent with Statistical Physics is that S/N tends to zero for N large when the temperature tends to the absolute zero. Notice that, in the case of the paramagnetic solid at zero external magnetic ﬁeld and very low temperature, one would expect to ﬁnd S/N = kB ln 2, in contradiction with the Nernst law. In fact, this spin system cannot be strictly in zero magnetic ﬁeld because of the nuclear magnetic interactions that are always present (but are not very eﬃcient at higher temperature where other phenomena imply much larger amounts of energy) : under the eﬀect of these

The Thermodynamical Potentials ; the Legendre Transformation

67

nuclear interactions the system is no longer at zero magnetic ﬁeld and becomes ordered when T tends to zero. The present sections have put in correspondence the laws of Thermodynamics and the statistical approach of chapters 1 and 2. Their conclusion is that the description based on Statistical Physics and its postulates is indeed equivalent to the three laws of Thermodynamics.

3.5

The Thermodynamical Potentials and their Diﬀerentials ; the Legendre Transformation

Now that we have identiﬁed the statistical entropy to the thermodynamical entropy, we can take advantage of our background in Thermodynamics to introduce the thermodynamical potential adapted to the considered physical situation, and relate it to the statistical description and to the physical parameters. The thermodynamical potential will be a function of the parameters determined by the experimental conditions and in equilibrium it will be extremum.

3.5.1

Isolated System

Here the entropy S is the thermodynamical potential to consider. During a spontaneous evolution toward equilibrium, S increases and reaches its maximum in equilibrium. In the microcanonical description, the entropy is related to the number W of microstates of energies lying in the range between E and E + dE by S = kB ln W

(3.23)

dS and dU are linked through the diﬀerential expressions of the ﬁrst and second laws of Thermodynamics : dS =

dU δW δQ = − T T T

(3.24)

so that 1 = T

Å

∂S ∂U

ã (3.25) N,Ω

68

Chapter 3. Thermodynamics and Statistical Physics

In the case of the ideal gas where δW = −P dΩ

dS =

(3.26)

dU P dΩ + T T

(3.27)

The diﬀerential expression (3.27) shows that S, calculated for a ﬁxed number N of particles, is a function of the variables U and Ω, both extensive. The pressure is deduced from the partial derivative of S versus the volume : Å ã ã Å P ∂S kB ∂ ln W (U ) = = (3.28) T ∂Ω N,U ∂Ω N,U Note : we suggested in the Appendix 3 of chapter 1 that, in the case of free particles, the number of microstates associated with the energy E was of the form W (E) ∝ ΩN χ(E)

(3.29)

so that P = T

Å

kB ∂ ln W (U ) ∂Ω

ã = N,U

kB N Ω

P Ω = N kB T

(3.30) (3.31)

Here we ﬁnd the ideal gas law, that will be again established in the next chapter. After having studied a given system, we now consider another similar system, with the same energy U and under the same volume Ω, which only diﬀers by its number of particles equal to N + dN . Its entropy is now S(U, Ω, N ) + dS, and we write µ dS = − dN T

(3.32)

This deﬁnes the chemical potential µ, related to the entropy variation of the system when an extra particle is added, at constant internal energy and volume. The chemical potential is an intensive parameter. For a classical ideal gas, if N increases, one expects S to increase, so that µ is negative. Then, including the dependence in the particles number, the diﬀerential of S for the ideal gas is written dS =

dU P dΩ µdN + − T T T

(3.33)

The Thermodynamical Potentials ; the Legendre Transformation

69

S is a function of extensive variables only ; the partial derivative of S versus N is equal to ã Å µ ∂S (3.34) =− ∂N Ω,U T We previously showed in § 2.5.1 that if the two systems are exchanging particles, in equilibrium ∂S = −kB α ∂N takes the same value in both systems.

(3.35)

These two partial derivatives of the entropy versus the number of particles can be identiﬁed, so that µ (3.36) α= kB T Besides, from the diﬀerential (3.33) of S expressed for the ideal gas one deduces the expression of the diﬀerential of its energy U : dU = −P dΩ + T dS + µdN

(3.37)

The last term can be interpreted as a “chemical work.”

3.5.2

System with a Fixed Number of Particles, in Contact with a Heat Reservoir at Temperature T

We are going to show that the corresponding thermodynamical potential is the Helmholtz free energy F , which is minimum in equilibrium and is related to the N -particle canonical partition function ZN . Indeed the general deﬁnition (3.16) of S applies, with now probabilities pi equal to pi =

1 exp(−βEi ) ZN

Then i

pi ln pi =

i

=−

Å pi

and β =

Ei − kB T

1 kB T

(3.38)

ã − ln ZN

U − ln ZN kB T

(3.39)

Consequently, in the canonical ensemble, U + kB ln ZN T U − T S = −kB T ln ZN S=

(3.40) (3.41)

70

Chapter 3. Thermodynamics and Statistical Physics

Now in Thermodynamics the free energy F is deﬁned by F = U − TS

(3.42)

whence we deduce the relation between the thermodynamical potential F , and the partition function ZN expressing the Statistical Physics properties : F = −kB T ln ZN

(3.43)

Let us write the diﬀerential of the thermodynamical potential F in the case where the inﬁnitesimal work is given by δW = − Xj dxj (see eq. 3.8) : j

dF = dU − T dS − SdT Xj dxj + δQ + µdN − T dS − SdT =−

(3.44)

j

Now δQ = T dS, so that dF = −

Xj dxj − SdT + µdN

(3.45)

j

The interpretation of (3.45) is that, in a speciﬁc inﬁnitesimal reversible process at constant temperature and ﬁxed number of particles, the free energy variation is equal to the work received by the system, whence the name of Helmholtz “free energy,” i.e., available energy, for this thermodynamical potential. For a gas, the inﬁnitesimal work is expressed as a function of pressure, so that dF = −P dΩ − SdT + µdN

(3.46)

The entropy, the pressure, and the chemical potential are obtained as partial derivatives F or of ln ZN : ã ã Å Å ∂(kB T ln ZN ) ∂F = (3.47) S=− ∂T N,Ω ∂T N,Ω ã ã Å Å ∂(kB T ln ZN ) ∂F = (3.48) P =− ∂Ω N,T ∂Ω N,T ã ã Å Å ∂(kB T ln ZN ) ∂F µ= =− (3.49) ∂N Ω,T ∂N Ω,T

3.5.3

System in Contact with Both a Heat Reservoir at Temperature T and a Particle Reservoir

It was shown in 2.5.1 that, during such an exchange of energy and particles, the parameters taking the same value in equilibrium in both the system and

The Thermodynamical Potentials ; the Legendre Transformation

71

the reservoir are kB β = 1/T and kB α = µ/T : two Lagrange parameters thus come into the probabilities, which do not only depend on the energy but also on the number of particles. Thus the probability of occurrence of a macroscopic state of energy Ei and number of particles Ni can be written (see § 2.5.3, eqs. (2.62) and (2.63)) : pi (Ei , Ni ) = so that

pi ln pi =

i

i

=−

1 exp(−βEi + αNi ) ZG

Å ã Ei + αNi − ln ZG pi − kB T

U + αN − ln ZG kB T

(3.50)

(3.51) (3.52)

Consequently, U − αkB N + kB ln ZG T U − T S − αkB T N = −kB T ln ZG S=

(3.53) (3.54)

where N is the average number of particles in the system under study. The thermodynamical potential introduced here is the grand potential A, deﬁned by A = U − T S − µN = F − µN

(3.55)

It is related to the grand canonical partition function through A = −kB T ln ZG

(3.56)

The diﬀerential of A is deduced from its deﬁnition : dA = dF − µdN − N dµ Xj dxj − SdT − N dµ =−

(3.57) (3.58)

j

Like the other thermodynamical potentials, A is extensive. The only extensive variable in A is the one in the inﬁnitesimal work. In the case of a gas, it is the volume so that A = −P Ω

(3.59)

The entropy S and the average number of particles N are obtained by calculating the partial derivatives of A. As a conclusion of this section on the thermodynamical potentials, remember that to a ﬁxed physical situation, corresponding to data on speciﬁc external parameters, is associated a particular thermodynamical potential, which is extremum in equilibrium.

72

3.5.4

Chapter 3. Thermodynamics and Statistical Physics

Transformation of Legendre ; Other Thermodynamical Potentials

In § 3.5.2 and § 3.5.3 we saw two diﬀerent physical situations, associated with two thermodynamical potentials with similar diﬀerentials, except for a change in one variable. The mathematical process that shifts from F (N ) to A(µ) is called “transformation of Legendre.” The general deﬁnition of such a transformation is the following : one goes from a function Φ(x1 , x2 , ...) to a function Γ = Φ(x1 , x2 , ...) − x1 y1 , with y1 =

∂Φ ∂x1

(3.60)

the new function Γ is considered as a function of the new variable y1 . Because of the change of variable and of function, the diﬀerentials of Φ and of Γ are respectively : ∂Φ ∂Φ dx1 + dx2 + . . . ∂x1 ∂x2 ∂Γ ∂Γ dy1 + dx2 + . . . dΓ = ∂y1 dx2

dΦ =

(3.61) (3.62)

Now ∂Γ ∂Γ ∂x1 = · = ∂y1 ∂x1 ∂y1

Å

∂y1 ∂Φ ∂Φ − − x1 ∂x1 ∂x1 ∂x1

ã ·

∂x1 = −x1 ∂y1

(3.63)

Consequently, dΓ = −x1 dy1 +

∂Φ dx2 + . . . ∂x2

(3.64)

The partial derivatives of Γ versus the variables x2 . . . are not modiﬁed with respect to those of Φ for the same variables. In this process one has shifted from the variables (x1 , x2 . . . xn ) to the variables (y1 , x2 . . . xn ). This is indeed the transformation that was performed for example, when shifting from U (S) to F (T ), the other variables Ω, N being conserved, or when shifting from F (N ) to A(µ), the variables Ω and T being unchanged. Other thermodynamical potentials can be deﬁned in a similar way, a speciﬁc potential corresponding to each physical situation. For example, a frequent situation in Chemistry is that of experiments performed at constant temperature and constant pressure (isothermal and isobaric condition) : the exchanges occur between the studied system A and the larger system A which acts as

The Thermodynamical Potentials ; the Legendre Transformation

73

an energy and volume reservoir. We saw that the equilibrium condition, expressing the maximum of entropy of the combined system with respect to exchanges of energy and volume, corresponds to the equality of the temperatures and the pressures in both systems. The corresponding thermodynamical potential is the Gibbs free enthalpy G, in which the number of particles is ﬁxed : G = U − T S + P Ω dG = ΩdP − SdT + µdN

(3.65)

This potential is deﬁned from the free energy F F = U − TS dF = −P dΩ − SdT + µdN

(3.66)

using a transformation of Legendre on the set of variables (volume Ω and pressure P ), that is G = F + P Ω

(3.67)

The chemical potential is interpreted here as the variation of the free enthalpy G when one particle is added to the system, at constant pressure and temperature. The free enthalpy, extensive like all thermodynamical potentials, depends of a single extensive variable, so that G = µN

(3.68)

The chemical potential µ is here the free enthalpy per particle. The free enthalpy is related to the partition function which includes all the possible values of energy and volume of the considered system (see § 2.6) : ZT,P =

e−βEi − γΩi

(3.69)

i

Note 1 : each time a transformation of Legendre is performed, an extensive variable is replaced by an intensive one. Now a thermodynamical potential has to be extensive, so that it should retain at least one extensive variable. Consequently, there is a limit to the number of such tranfomations yielding a new thermodynamical potential. Note 2 : Appendix 4.1 presents a practical application of thermodynamical potentials to the study of chemical reactions.

Summary of Chapter 3 We have expressed the laws of Thermodynamics in the language of Statistical Physics. – Zeroth law : For two systems A and A in thermal equilibrium ã Å 1 ∂S 1 = kB β = = kB β = T ∂E N,Ω T – First law : The diﬀerential of the average energy of a system : dU = d pi Ei = pi dEi + Ei dpi i

i

i

is split into two terms. The inﬁnitesimal work is related to the modiﬁcation of the energy states of the system due to the change of external parameters (volume, etc.) δW = pi dEi i

The inﬁnitesimal heat exchanged δQ is associated with variations of probabilities of occurrence of the microstates Ei dpi δQ = i

– Second law : We have identiﬁed the thermodynamical entropy with the statistical entropy as introduced in the ﬁrst chapters of this book and have shown that 1 kB β = . T 75

76

Summary of Chapter 3

In Statistical Physics, an irreversible process corresponds to an increase of the number of accessible microstates for the considered system, and consequently to an increase of disorder and of entropy. – Third law : The entropy per particle S/N vanishes when the temperature tends to the absolute zero. – The adapted thermodynamical potential depends on the considered physical situation, it is extremum in equilibrium. It is related to the microscopic parameters of Statistical Physics through the partition function ; its partial derivatives give access to the measurable physical parameters. Using a transformation of Legendre, one gets a new thermodynamical potential, with one new variable (see the table next page).

Parameters

Thermodynamical

dS =

P dU µ dΩ + − dN T T T 1 ∂S ... = T ∂U Ω,N

S = kB ln W

Thermodynamical Potential

microstates

S=−

∂F ∂T

e−βEn

Ω,N

...

dF = −P dΩ − SdT + µdN

F = −kB T ln ZN F = U − TS

ZN (β) =

W microscopic conﬁgurations at E

Statistical Description

1 kB T

exchange of energy with a heat reservoir at ﬁxed N

β=

isolated system E to δE

Physical Conditions

Canonical

Lagrange Parameter(s)

Microcanonical

Ensemble

eαN ZN (β)

N = −

Å

∂A ∂µ

Ω,T

ã

(average value)

dA = −P dΩ − SdT − N dµ

A = −kB T ln ZG A = U − T S − µN

N

1 µ , α = βµ = kB T kB T

ZG (α, β) =

β=

exchange of energy with a heat reservoir and of particles with a particle reservoir

Grand Canonical

Summary of Chapter 3 77

Chapter 4

The Ideal Gas 4.1

Introduction

After having stated the general methods in Statistical Physics and bridged its concepts with your previous background in Thermodynamics, we are now going to apply our know-how to the molecules of a gas, with no mutual interaction. The properties of such a gas, called “ideal gas” (“perfect gas” in French) are discussed in Thermodynamics courses. Here the contribution of Statistical Physics will be evidenced : starting from a microscopic description in Classical Mechanics, it allows to explain the observed macroscopic properties. A few notions of kinetic theory of gases will ﬁrst be introduced in § 4.2 and § 4.3 we will discuss to what extent classical Statistical Physics, as used up to this point in this book, can apply to free particles. This chapter is “doubly classical” since, like in the three preceding chapters, Classical Statistics is used, but now for particles with an evolution described by Classical Mechanics : both the states themselves and their occurrence are classical. The canonical ensemble is chosen in § 4.4 to solve the statistical properties of the ideal gas and physical consequences are deduced ; in the Appendices, all that we learnt in this chapter and the preceding ones is applied in a practical problem : the interpretation of chemical reactions in gas phase is done in Thermodynamics (Appendix 4.1) and in Statistical Physics (Appendix 4.2).

79

80

4.2

Chapter 4. The Ideal Gas

Kinetic Approach

This section, which has no relation with Quantum Mechanics or Statistical Physics in equilibrium, will allow us to obtain insight into the simplest physical parameters describing the motion of molecules in a gas, and to introduce some orders of magnitude. If necessary, we will refer to the Thermodynamics courses on the ideal gas.

4.2.1

Scattering Cross Section, Mean Free Path

Consider a can ﬁlled with a gas, in the standard conditions of temperature and pressure (T = 273 K, P = 1 atmosphere). Each molecule has a radius r, deﬁned by the average extension of the valence orbitals, i.e., a fraction of a nanometer. If in its motion a molecule comes too close to another one, there will be a collision. One deﬁnes the scattering cross section which is the circle of radius r around a molecule, inside which no other molecule can penetrate.

l=

ντ

r

d Fig. 4.1: Deﬁnitions of the scattering cross section and of the mean free path. The mean distance between two collisions l, or mean free path, is deduced from the mean volume per particle Ω/N and the scattering cross section by l(πr2 ) =

Ω = d3 N

(4.1)

where d is the average distance between particles (Fig. 4.1). From l, one deduces the mean time τ between two collisions, or collision time, by writing l = ντ where ν is the mean velocity at temperature T , of the order of the justiﬁcation in § 4.4.2).

(4.2) kB T /m (see

Kinetic Approach

81

For nitrogen molecules with r = 0.42 nm, at 273 K and under a pressure of one atmosphere one ﬁnds l ≈ 70 nm, ν ≈ 300 m.sec−1 , τ ≈ 2 × 10−10 sec. The length l is to be compared with the average distance between particles d = 3.3 nm. These values show that in the gas collisions are very frequent at the time scale of a measure, which ensures that equilibrium is rapidly reached after a perturbation of the gas ; at the molecular scale distances covered between collisions are very large (of the order of a hundred times the size of a molecule), thus the motion of each molecule mostly occurs as if it was alone, that is, free.

4.2.2

Kinetic Calculation of the Pressure

Let us now recall the calculation, from kinetic arguments, of the pressure of a gas, i.e., of the force exerted on a wall of unit surface by the molecules. The aim is to evaluate the total momentum transferred to this wall by the molecules impinging on it during the time interval ∆t : indeed in this chapter Classical Mechanics does apply, so that the total force on the wall is given by the Fundamental Equation of Dynamics : Ftotal =

∆ pi i

(4.3)

∆t

Here the momentum brought by particle i is ∆ pi . The resulting force is normal since the wall is at rest, it is related to the pressure P and to the considered wall surface S through |Ftotal | = P S

(4.4)

One will now evaluate the total momentum transferred to the surface, which is the diﬀerence between the total momentum of the particles which reach the wall and that of the particles which leave it (Fig. 4.2). pi f (r, p i , t) The incident particles, with pz > 0, bring a total momentum i

to the wall, chosen as xOy plane. The probability density in equilibrium f (r, pi , t) is assumed to be stationary, homogeneous, and isotropic in space (hypotheses stated by James Clerk Maxwell in 1859). It thus reduces to an even function f ( p) versus the three components of the momentum and satisﬁes to (4.5) f ( p)d3 p = 1 Consider the particles, of momentum equal to p to d3 p, incident on the surface S during the time interval ∆t. They are inside an oblique cylinder along the

82

Chapter 4. The Ideal Gas

pz ∆t m S

p

O

z

p

Fig. 4.2 : Collisions of molecules on a wall. Spz pz ∆t and volume ∆t (the direction z is m m chosen along the normal to the surface). For a particles density ρ = N/Ω, the number of incident particles is

direction of p, of basis S, height

Spz ∆t ρf ( p)d3 p m

(4.6)

Each of them transfers its momentum p to the wall. The total momentum brought by the incident particles is thus equal to S ∆ pi = ∆tρ pz pf ( p)d3 p (4.7) m p >0 z inc This sum is restricted to the particles traveling toward the surface, of momentum pz > 0. A similar calculation leads to the total momentum of the particles which leave the surface, or in other words “were reﬂected,” by the wall. The volume (−pz ) is S ∆t, the total momentum taken from the wall is equal to m S ∆ pi = ∆tρ (−pz ) pf ( p)d3 p (4.8) m pz 0. Indeed one will always perform an integration by parts of a seconddegree term in energy in the numerator to express the corresponding factor of the denominator partition function, which will provide the factor kB T /2. But this property is not valid for a polynomial of degree diﬀerent than two, for example for an anharmonic term in x3i . This result constitutes the “theorem of energy equipartition”, established by James C. Maxwell as early as 1860 : For a system in canonical equilibrium at temperature T , the average value of the energy per particle and per quadratic degree of 1 freedom is equal to kB T . 2 We have already used this property several times, in particular in § 4.3 to ﬁnd the root-mean-square momentum of the ideal gas at temperature T . It is also useful to determine the thermal ﬂuctuations of an oscillation, providing 1 1 a potential energy in kx2 for a spring, in Cϑ2 for a torsion pendulum, 2 2 and more generally to calculate the average value at temperature T of a hamiltonian of the form h= ai x2i + bj p2j (4.33) i

j

where ai is independent of xi , bj does not depend of pj , and all the coeﬃcients ai and bj are positive (which allows to use this property for an energy development around a stable equilibrium state). A case of application of the energy equipartition theorem is the speciﬁc heat at constant volume associated to U . We recall its expression ã Å ∂U Cv = (4.34) ∂T Ω,N

Classical Statistics Treatment of the Ideal Gas

89

In the case of a monoatomic ideal gas, Cv only contains the contribution of 3 the three translation degrees of freedom and is thus equal to N kB . 2

4.4.3

Free Energy ; Physical Parameters (P, S, µ)

We saw in chapter 3 that the canonical partition function Zc is related to the free energy through F = U − T S = −kB T ln Zc

(4.35)

In the case of an ideal gas, the diﬀerential of F is dF = −P dΩ − SdT + µdN

(4.36)

The volume Ω and the number of particles N are extensive variables, the temperature T is intensive. The chemical potential, already introduced in § 3.5.3, is here the variation of free energy of the system at ﬁxed temperature and volume when one particle is added. Substituting the expression of Zc obtained in § 4.4.1, one obtains : Å ã Ω F = −kB T N ln 3 (2πmkB T )3/2 + ln CN h ã Å Ω 1 3/2 F = −N kB T ln ln C (2πmk T ) + ln N + B N N h3 N

(4.37) (4.38)

For the potential F to be extensive, the quantity between brackets must be Ω should appear in the ﬁrst term and intensive. This dictates that the ratio N that the remaining terms should be constant : ln N +

1 ln CN = constant N

(4.39)

i.e., ln CN = −N ln N + constant × N

(4.40)

From now on, we will take CN =

1 , i.e. ln CN = −N ln N + N + O(ln N ) N!

(4.41)

so that the constant in Eq. (4.39) is equal to unity. This choice will be justiﬁed in § 4.4.4 by a physical argument, the Gibbs paradox, and will be demonstrated in § 6.7. Then ï Å Å ã ò ã Ω eΩ 3/2 (2πmk T ) T ln +1 = −N k (4.42) F = −N kB T ln B B N h3 N λ3th

90

Chapter 4. The Ideal Gas

One deduces the pressure, the entropy and the chemical potential of the assembly of N particles : Å ã ∂F N kB T (4.43) = P =− ∂Ω N,T Ω ã ã Å Å 5 ∂F Ω + N kB S=− = N kB ln (4.44) ∂T N,Ω N λ3th 2 ã ã Å Å Ω ∂F (4.45) = −kB T ln µ= ∂N T,Ω N λ3th One identiﬁes the ﬁrst of these relations as the ideal gas law, that we now just re-demonstrated using Statistical Physics. The expression obtained for S(Ω, N, T ) is extensive but does not satisfy the Nernst principle, since it does not vanish when the temperature tends to zero : the entropy would become very negative if the de Broglie thermal wavelength was becoming too large with respect to the mean distance between particles. This conﬁrms what we already expressed in § 4.3, that is, the ideal gas model is no longer valid at very low temperature. In the validity limit of the ideal gas model, λth d is equivalent to Ω/N λ3th very large, thus to µ very negative (µ → −∞). Remember that the measure of the classical N -particle phase space must be N 1 d3ri d3 pi taken equal to to ensure the extensivity of the thermodynaN ! i=1 h3 mical potentials F and S. This will be justiﬁed in chapter 6 : the factor CN expresses the common classical limit of both Quantum Statistics.

4.4.4

Gibbs Paradox

At the end of the 19th century, John Willard Gibbs used the following physical argument to deduce the requirement of the constant CN = N1 ! in the expression of the inﬁnitesimal volume of the phase space : consider two containers of volumes Ω1 and Ω2 , containing under the same density and at the same temperature T , respectively, N1 and N2 molecules of gas (Fig. 4.3). One then opens the partition which separates the two containers, the gases mix and the ﬁnal total entropy is compared to the initial total one. If the gases in the two containers were of diﬀerent chemical natures, the disorder and thus the entropy increase. On the other hand, if both gases are of the same nature and the molecules indistinguishable, the entropy is unchanged in the process. The detailed calculation of the diﬀerent terms in the total entropy, before

Conclusion

91

N1 et N2 of the same nature Ω 1 + Ω2 N1 + N2 T Ω1 N1 S1 T

Ω2 N2 S2 T

N1 N2 = Ω1 Ω2

S = S1 + S2 N1 et N2 of diﬀerent natures Ω 1 + Ω2 N1 , N2 T S > S1 + S2

Fig. 4.3: Mixture of two gases at the same temperature and with the same density. and after opening the partition, is simple and uses the partial derivative with respect of temperature of expression (4.38) of the free energy, including the constant CN : if one assumes that CN = 1, there will be a variation of the total entropy even for a mixture of molecules of the same nature. When one introduces the factor CN = 1/N !, the result corresponds to the physical prediction. The choice of this factor is not surprising once one considers that molecules of the same chemical nature are indistinguishable : for a system of N indistinguishable particles, there are N ! possible permutations of these particles among their one-particle states, providing the same macroscopic state. In order to account only once for this state in the enumeration that will lead to the physical predictions, among others of the entropy, it is natural to choose the value CN = 1/N !.

4.5

Conclusion

Using Statistical Physics, we have been able to ﬁnd again the well-known results of Thermodynamics on the ideal gas and to justify them. It seems interesting, now that we come to the end of the study of Classical Statistics, to present the respective contributions of Thermodynamics (Appendix 4.1) and of Statistical Physics (Appendix 4.2) on a practical example. This will be done on the analysis of the chemical reactions. In the Statistical Physics part, one will show in details the various microscopic terms of the hamiltonian of a gas of diatomic molecules ; this allows one to

92

Chapter 4. The Ideal Gas

understand the origin of a law very commonly used in Chemistry, that is, the law of mass action. Enjoy your reading ! Reading these two Appendixes is not a prerequisite to understanding the next chapters.

Summary of Chapter 4 First we have presented some orders of magnitude, relative to nitrogen molecules in standard conditions of temperature and pressure. They justify that the hypotheses of the ideal gas, of molecules with no mutual interaction, is reasonable : indeed the distances traveled by a molecule between collisions are very large with respect to a molecule radius and the characteristic time to reach equilibrium is very short as compared to a typical time of measurement. Then the kinetic calculation of a gas pressure was recalled. The domain of validity of the ideal gas model, in the framework of both Classical Mechanics and Classical Statistics, is the range of high temperatures and low densities : the criterion to be fulﬁlled is Å ã1/3 Å ã1/3 Ω h Ω

d= 2πmkB T , or λth = √ h N N 2πmkB T where λth is the thermal de Broglie wavelength and d the mean distance between particles. The statistical treatment of the ideal gas is done in this chapter in the canonical ensemble, the N -particle canonical partition function is given by ïÅ ã òN Ω 3/2 Zc = CN (2πmk T ) B h3 To ensure that the free energy and the entropy are extensive and to solve the Gibbs paradox, one must take CN = 1/N !. The average value at temperature T of the energy associated to a quadratic 1 degree of freedom is equal to kB T (theorem of the equipartition of energy). 2 This applies in particular to the kinetic energy terms for each coordinate. The chemical potential of the ideal gas is very negative ; its entropy does not satisfy to the third law of Thermodynamics (but the ideal gas model is no longer valid in the limit of very low temperatures). 93

Appendix 4.1 Thermodynamics and Chemical Reactions We will consider here macroscopic systems composed of several types of molecules or chemical species, possibly in several phases : one can study a chemical reaction between gases, or between gases and solids, etc. The ﬁrst question that will be raised is “What is the spontaneous evolution of such systems ?” ; then the conditions of chemical equilibrium will be expressed. These types of analyses, only based on Thermodynamics, were already introduced in your previous studies. Now Appendix 4.2 will evidence the contribution of Statistical Physics. 1. Spontaneous Evolution In the case of a thermally isolated system, the second law of Thermodynamics indicates that a spontaneous evolution will take place in a direction corresponding to an increase in entropy S, S being maximum in equilibrium.

Q A A T0

Atot Fig. 4.4: The system A is in thermal contact with the heat reservoir A , with which it exchanges the heat Q. The combined system is isolated. 95

96

Appendix 4.1

Chemical reactions are mostly done at constant temperature : one then has to consider the case of a system A in thermal contact with a heat reservoir A at temperature T0 (Fig. 4.4). The combined system Atot , made of A and A is an isolated system, the entropy Stot of which increases in a spontaneous process : ∆Stot ≥ 0

(4.46)

The entropy variation of the combined system ∆Stot = ∆S + ∆S

(4.47)

can be expressed as a function of parameters of the system A alone. Indeed, if the amount of heat Q is given to A by the heat reservoir, the reservoir absorbs −Q and its entropy variation is equal to ∆S = −Q/T0 : the transfer of Q does not produce a variation of T0 and one can imagine a reversible sequence of isothermal transformations at T0 producing this heat exchange. Besides, the internal energy variation in A, i.e., ∆U , is, from the ﬁrst law of Thermodynamics, ∆U = W + Q

(4.48)

where W is the work operated in the considered transformation. Consequently, ∆Stot = ∆S −

(∆U − W ) −∆F0 + W = T0 T0

(4.49)

where, by deﬁnition, F0 = U − T0 S, which is identical to the Helmholtz free energy F = U − T S of system A if the temperature of A is that of the heat reservoir (in an oﬀ-equilibrium situation T diﬀers from T0 and F is not identical to F ). From the condition of spontaneous evolution one deduces that −∆F0 ≥ (−W )

(4.50)

which expresses that the maximum work that may be delivered by A in a reversible process is (−∆F0 ), whence the name of “free energy.” If the volume of system A is maintained constant and there is no other type of work, W = 0, so that the condition for spontaneous evolution is then given by ∆F0 ≤ 0

(4.51)

Appendix 4.1

97

This condition for a system in contact with a heat reservoir replaces the condition ∆Stot ≥ 0 for an isolated system. A more realistic situation is the one in which system A is kept at constant temperature and constant pressure. This is the case in laboratory experiments when a chemical reaction is performed at ﬁxed temperature T0 and under the atmospheric pressure P0 : system A is exchanging heat and work with a very large reservoir A , which can thus give energy to A without any modiﬁcation of T0 and also exchange a volume ∆Ω with A without any change of its pressure P0 (Fig. 4.5). A T0 P0

A Q T ∆Ω P

Atot Fig. 4.5: System A is in contact with the larger system A , with which it is exchanging the amount of heat Q and the volume ∆Ω. Similarly, one considers that the combined system made of A and A is isolated, so that in a spontaneous process ∆Stot = ∆S + ∆S ≥ 0

(4.52)

If A absorbs Q from A , ∆S = −Q/T0 , where T0 is again the temperature of A . The variation of internal energy of A includes the received heat Q, the mechanical work −P0 ∆Ω received by A from A , and also W ∗ , work of any other origin (electrical for example) that A may have received : ∆U = Q − P0 ∆Ω + W ∗

(4.53)

Consequently, ∆Stot = ∆S − i.e.,

∆Stot =

Q T0 ∆S − (∆U + P0 ∆Ω − W ∗ ) = T0 T0

−∆G0 + W ∗ T0

(4.54) (4.55)

98

Appendix 4.1

Here we deﬁned G0 = U + P0 Ω − T0 S, which identiﬁes to the Gibbs free enthalpy of system A, that is, G = U + P Ω − T S, when P = P0 and T = T0 . A spontaneous evolution will be expressed through the condition −∆G0 + W ∗ ≥ 0

(4.56)

When all the external parameters of system A, except its volume, are kept ﬁxed, W ∗ is null and the condition of spontaneous evolution is then ∆G0 ≤ 0

(4.57)

Consequently, when a system is in contact with a reservoir so that it remains at ﬁxed temperature and pressure and can exchange work only with this reservoir, the stable equilibrium is characterized by the condition of minimum of G0 . 2. Chemical Equilibrium, Law of Mass Action : Thermodynamical Approach Consider a homogeneous system (for example in gaseous phase) with several types of molecules : here we will study the dissociation reaction of iodine vapor at temperatures in the 1000 K range : I2 2I In this reaction the number of iodine atoms is conserved between the species corresponding to the two members of the chemical equation, which can also be expressed in the form 2I − I2 = 0 the species disappearing in the reaction carrying a negative sign, the one appearing a positive one. In the same way, the most general chemical equilibrium will be written n

bi Ni = 0

(4.58)

i=1

The bi are the integral coeﬃcients of the balanced reaction for the Bi molecules. The number Ni of molecules of Bi type in the system is changed if the equilibrium is displaced but the total number of atoms of each species in the system is conserved. The variations of the numbers Ni are thus proportional to the coeﬃcients bi of the chemical equilibrium dNi = bi dλ

for any i

where dλ is a proportionality constant. dNi is positive for the molecules appearing, negative for those vanishing in the reaction.

Appendix 4.1

99

Consider a chemical equilibrium at ﬁxed temperature T and pressure P , thus characterized by dG(T, P ) = 0. In such conditions the equilibrium is expressed by : dG = At ﬁxed T and P, dG =

n i=1 n

dGi =

n

(−Si dT + Ωi dP + µi dNi )

i=1

(4.59)

µi dNi = 0

i=1

Å

ã ∂G Here µi = is the chemical potential of species Bi . Replacing dNi ∂Ni T,P by its expression versus λ, one obtains the relation characterizing the equilibrium : n

bi µi = 0

(4.60)

i=1

The chemical potentials depend on the experimental conditions, in particular on the concentrations of the diﬀerent species. Assume that these reagents are ideal gases and look for the dependence of µi on the partial pressure Pi of one of these gases. Since Gi is extensive and depends on a single extensive parameter Ni , the number of molecules of the species i, then one must have Gi = µi Ni

(4.61)

The dependence in temperature and pressure of µi is thus that of Gi . For an ideal gas, at ﬁxed temperature, Gi = Ui + Pi Ω − T Si

(4.62)

introducing the partial pressure Pi of species i, such that Pi Ω = Ni kB T . Consequently, dGi = −Si dT + ΩdPi + µi dNi Å

∂Gi ∂Pi

ã

Å = Ni

T,Ni

∂µi ∂Pi

ã =Ω= T,Ni

Ni kB T Pi

(4.63)

(4.64)

i.e., by integrating (4.64), µi = µ0i (T ) + kB T ln

Pi P0

(4.65)

100

Appendix 4.1

where µ0i (T ) is the value of the chemical potential of species i at temperature T and under the standard pressure P0 . The general condition of chemical equilibrium is thus given by Å ãbi n n Pi bi µ0i + kB T ln =0 P0 i=1 i=1

(4.66)

One writes ∆G0m = bi µ0i N i=1 n

(4.67)

∆G0m , the molar free enthalpy at the temperature T of the reaction, is the diﬀerence between the molar free enthalpy of the synthesized products and that of the used reagents, N is the Avogadro number. One also deﬁnes Kp =

ã n Å Pi bi i=1

P0

(4.68)

For example, in the dissociation equilibrium of iodine molecules, Kp =

(PI /P0 )2 (PI2 )/P0

(4.69)

One recognizes in Kp the constant of mass action. The above equation gives 0

Kp = e−∆Gm /RT

(4.70)

where R = N kB is the ideal gases constant, R = 8.31 J.K−1 . From the tables of free enthalpy values deduced from experiments, one can deduce the value of Kp . From the experimental study of a chemical equilibrium at temperature T , one can also deduce the value of Kp , and thus of ∆G0m , at temperature T : if the temperature of the experiment is modiﬁed from T to T , the partial pressures are changed, Kp (T ) becomes Kp (T ), such that ln Kp (T ) = −∆G0m (T )/RT

(4.71)

By deﬁnition, at a given T , ∆G0m = ∆H 0 − T ∆S 0 , where ∆H 0 is the molar enthalpy of the reaction and ∆S 0 its entropy variation. If T and T are close enough, one makes the approximation that neither ∆H 0 nor ∆S 0 noticeably vary between these temperatures. Then Å ã 1 1 (4.72) − ln Kp (T ) − ln Kp (T ) = −∆H 0 RT RT

Appendix 4.1

101

i.e., also ∆H 0 d ln Kp (T ) = dT RT 2

(4.73)

From the enthalpy ∆H 0 , one can predict the direction of displacement of the equilibrium when the temperature increases : if ∆H 0 > 0, that is, the enthalpy increases in the reaction and the reaction is endothermal, the equilibrium is thus displaced toward the right side when T increases. This is in agreement with the principle of Le Châtelier : If a system is in stable equilibrium, any spontaneous change of its parameters must lead to processes which tend to restore the equilibrium. As an application, the experimental data below on the dissociation of iodine allow to deduce the enthalpy of this reaction in the vicinity of 1000 K. For a ﬁxed total volume of 750.0 cm3 , in which a total number n of molecules I2 was introduced at temperature T , the equilibrium pressure is P : T (K) P (10−2 atm) n(10−4 moles)

973 6.24 5.41

1 073 7.50 5.38

1 173 9.18 5.33

As an exercise, one can calculate Kp at each temperature and verify that ∆H 0 = 157 kJ.mole−1 . This reaction is endothermal. The example presented here suggests the rich domain, but also the limitations of the applications of Thermodynamics in Chemistry : Thermodynamics does not reduce to a game between partial derivatives, as some students may believe ! It allows one to predict the evolution of new reactions, from tables deduced from measures. However, only the contribution of Statistical Physics allows us to predict these evolutions from the sole spectroscopic data and microscopic models. It is this latter approach that will be followed now. Thanks to the understanding acquired in this ﬁrst part of the Statistical Physics course, we will be able to deduce the constant of mass action from the quantum properties of the chemical species of the reaction.

Appendix 4.2 Statistical Physics and Chemical Reactions The same example of dissociation of the iodine molecule at 1000 K, in gas phase, will now be treated by applying the general methods of Statistical Physics. The calculations will be completely developed, which may be regarded as tedious but yet they are instructive ! 1. Energy States : The Quantum Mechanics Problem Iodine I is a monoatomic gas, its only degrees of freedom are those of translation. The electronic structure of its fundamental state is 2 P3/2 (notation 2S+1 LJ : orbital state = 1 whence the notation P ; spin 1/2, i.e., 2S + 1 = 2 ; total kinetic moment equal to J = 3/2, therefore the degeneracy 4). Iodine has a nuclear moment J=5/2. The hamiltonian of the iodine atom in the gas phase is therefore given by HI = HI transl + HI elec + HI nucl

(4.74)

The molecule I2 has six degrees of freedom, which can be decomposed into three degrees of translation of its center of mass, two of rotation and one of vibration : HI2 = HI2 transl + Hrot + Hvibr + HI2 elec + HI2 nucl

(4.75)

Since this molecule is symmetrical, one should account for the fact that the two iodine atoms are indistinguishable inside the molecule and thus the physics of the system is invariant in a rotation of 180 degrees around an axis perpendicular to the bond in its middle. The fundamental electronic state of molecular iodine is not degenerate. It lies lower than the fundamental electronic state of atomic iodine. The dissociation energy, necessary at zero temperature to transform a molecule into two iodine 103

104

Appendix 4.2

atoms, is equal to Ed = 1.542 eV, which corresponds to 17,900 K : one should bring Ed /2 per atom to move from the state I2 to the state I. The nuclear spin of iodine has the value J = 5/2, both under the atomic and molecular forms ; one assumes that no magnetic ﬁeld is applied. The splittings between energy levels of I or I2 are known from spectroscopic measures : Raman or infrared techniques give access to the distances between rotation or vibration levels ; the distance between electronic levels are generally in the visible (from 1.5 eV to 3 eV) or the ultraviolet range. One now expresses the eigenvalues of the diﬀerent terms of HI and HI2 . a. Translation States One has to ﬁnd the hamiltonian eigenstates : in the case of the iodine atom of mass mI , Htr =

pˆ2 + V (r) 2mI

(4.76)

where V (r) is the conﬁnement potential limiting the particle presence to a rectangular box of volume Ω = Lx Ly Lz . √ The eigenstates are plane waves ψ = (exp ik · r)/ Ω ; the k values are quantized and for the Born-Von Kármán periodic limit conditions [ψ(x+Lx ) = ψ(x)] (see § 6.4.2) the k components satisfy : kx Lx = nx 2π ; ky Ly = ny 2π ; kz Lz = nz 2π with nx , ny , nz positive, negative or null integers. The corresponding energies are : 2 2 k 2mI ñÅ ã2 Å ã2 Å ã2 ô nx h2 ny nz = + + 2mI Lx Ly Lz

ε I tr =

(4.77)

The translation energies εI2 tr of the molecule I2 are obtained by replacing mI by mI2 in the above expression of εItransl . All these states are extremely close : if Ω is of the order of 1 cm3 and the dimensions of the order of 1 cm, the energy of the state nx = 1, ny = nz = 0 is given by, for I (mI = 127 g) Å ã2 a0 1 mH × (0.53 × 10−8 )2 = 3 × 10−18 eV ! · = 13.6 × 13.6 eV × mI Lx 127

Appendix 4.2

105

The distances between energy levels are in the same range, thus very small with respect to kB T , for T = 1000K. This discussion implies that the ideal gas model is valid at room temperature for I and I2 (still more valid than for N2 , since these molecules are heavier). The partition functions for their translation states will take the form (4.26)(4.27).

b. Rotation States The iodine molecule is linear (Fig. 4.6), its inertia momentum around an Å ã2 Re axis normal to the bond is I = 2m = µRe2 with µ = mI /2 and Re 2 the equilibrium distance between the two nuclei. Since mI = (127/N )g this momentum is particularly large at the atomic scale. The rotation hamiltonian is 1 ˆ2 J 2I and shows that the problem is isotropic in the center of mass frame. Hrot =

(4.78)

Re

mI

mI

Fig. 4.6 : Structure of the I2 molecule. Its eigenvalues are : εrot (J) =

2 J(J + 1) = hcBJ(J + 1), 2I

where

B=

4πcI

is given in cm−1 (4.79)

(one recalls the deﬁnition hν = hc/λ = hcσ ; the unit cm−1 for σ is particularly adapted to infrared spectroscopy). The selection rules ∆J = ±1 for allowed electric dipolar transitions between rotation levels correspond to εrot (J + 1) − εrot (J) = hc · 2B(J + 1). For iodine B = 0.0373 cm−1 , 2hcB/kB = 0.107 K, to be compared to HC where B = 10.6 cm−1 , 2hcB/kB = 30.5 K. The rotation levels are thus still very close at the thermal energy scale.

106

Appendix 4.2

c. Vibration States In a ﬁrst approximation the vibration hamiltonian of the molecule I2 can be described as a harmonic oscillator hamiltonian, with a parabolic potential well centered at the equilibrium distance Re between the two nuclei, that is, V (R) = 12 K(R−Re )2 (Fig. 4.7). (In this ﬁgure a realistic potential is sketched in full line ; the dotted line represent its approximated harmonic potential.) ε

3 2 1 n=0 Re

R

Fig. 4.7 : Potential well associated to the vibration of the molecule I2 .

Hvibr =

1 pˆ2R + K(R − Re )2 2µ 2

(4.80)

∂ with pˆR = −i , µ being the reduced mass of the system. In this approxi∂R mation, the vibration energies are given by ã Å K 1 εvibr (n) = n + ω with ω = (4.81) 2 µ The transitions ∆ε = ω = hcσv are in the infrared domain. For I2 σv = 214.36 cm−1 , which corresponds to λ = 46.7µm or hcσv /kB = 308 K (for F2 the corresponding temperature is 1280 K).

d. Electronic States The splitting between fundamental and excited states, for both I and I2 , is of the order of a few eV, corresponding to more than 10,000 K. The origin of the electronic energies (Fig. 4.8) is taken at the I2 fundamental state, so that the energy of the I fundamental state is EI , with 2EI = Ed = 1.542 eV, where Ed is the dissociation energy.

Appendix 4.2

107

εe2 εe1 εe1 g=4

εe0 EI

g=1 I2

εe0

I

Fig. 4.8 : Electronic states of I2 and I. e. Iodine Nuclear Spin It is J = 5/2 on both species I and I2 . 2. Description of the Statistical State of the System : Calculation of the Partition Functions It was stated in Ch.2 that for a macroscopic system the predictions for the experimental measurements of physical parameters were the same whatever the chosen statistical ensemble. Here we will work in the canonical ensemble, where the calculations are simpler. At temperature T , the number of iodine molecules is NI2 , the number of atoms produced by the dissociation of the molecules NI . The partition function for a single iodine molecule is written zI2 , that of an isolated atom zI . Because of the particles indistinguishability, the total canonical partition function Zc is given by : Zc =

1 1 (zI )NI2 (zI )NI NI2 ! 2 NI !

(4.82)

It is related to the free energy F of the system F = −kB T ln Zc

(4.83)

the partial derivatives of which, with respect to the numbers of particles, are the chemical potentials : ã Å Å ã ∂F zI2 zI µI = −kB T ln (4.84) = −kB T ln µI2 = ∂NI2 NI2 NI Since the total energy of an iodine atom or molecule is a sum of contributions, the partition function of a corresponding atom or a molecule is a product of

108

Appendix 4.2

factors : thus for an I2 molecule εI2 = εI2 tr + εrot (J) + εvibr (n) + εel + εnucl

(4.85)

zI2 = zI2 tr · zrot · zvibr · zel · znucl

(4.86)

with :

2 1 3 r d3 p e−βp /2mI2 d 3 h = gJ e−βεJ = (2J + 1)e−βhcBJ(J+1)

zI2 tr = zrot

J

zvibr =

(4.88)

J

gv e−βεv =

v

zel =

(4.87)

1

e−βω(n+ 2 )

(4.89)

n

ge e−βεe

(4.90)

e

znucl = (2J + 1)2 for I2 ,

znucl = 2J + 1 for I

(4.91)

The factors gr , gv , ge express the degeneracies of the rotation, vibration and electronic energy levels. The value of the nuclear spin is J , there are two nuclei in I2 . One will ﬁrst evaluate the various partition functions, noting that a large value of the partition function is related to a large number of accessible states. The examined chemical potentials will be sums of terms, associated with the zI zI diﬀerent z factors. Note that one has to calculate 2 (or ) (see (4.84)), NI2 NI thus to divide a single factor of zI2 by NI2 : we will associate the factor NI2 to the sole translation factor, then it no longer comes into the partition functions for the other degrees of freedom. (The indiscernability factor is thus included in the translation degree of freedom.) Then : ln

ztr I2 zI2 = ln + ln zrot + ln zvibr + ln zel + ln znucl NI2 NI2

(4.92)

a. Translation The calculation of ztr I was already done for the ideal gas : Å

ztr I

2πmI kB T =Ω· h2

ã3/2 =

Ω NI kB T 1 = · (λth I )3 Pi (λth I )3

(4.93)

Pi is the partial pressure, such that Pi Ω = NI kB T and λth I is the thermal de Broglie wavelength at T for I. Since we are in the classical framework, ztr I

Appendix 4.2

109

must be very large. Numerically λth , expressed in picometers (10−12 m), is given by 1749 · 10−12 m λth = (TK × mg·mole−1 )1/2 At 1000 K, for the iodine atom mI = 127g,

λth I = 4, 91 10−12 m.

In the same condition, for the I2 molecule,

λth I2 = 3.47 10−12 m.

For a partial pressure equal to the atmospheric pressure and a temperature T = 1000 K, one obtains zI tr = 1.15 × 109 NI zI2 tr = 3.26 × 109 NI2 These numbers are huge !

b. Rotation The factor 2J + 1 is the degeneracy of the energy level hcBJ(J + 1). Besides, the linear molecule I2 is symmetrical, so that after a 180 degree rotation one gets a new state, indistinguishable from the initial one (this would not be the case for HC for example). One thus has to divide the above expression (4.88) by two. This is equivalent to saying that the wave function symmetry allows to keep only half of the quantum states. Since the energy distance between rotation levels is small with respect to kB T , one can replace the discrete sum by an integral 1 1 ∞ (2J + 1)e−βhcBJ(J+1) (2J + 1)e−βhcBJ(J+1) dJ 2 2 0 J (4.94) 1 1 kB T = 2 βhcB 2hcB This approximation implies that zrot is a large number. Numerically, the contribution of zrot , equal to 1/2βhcB, is the ratio 1000 K/0.107 K=9346.

c. Vibration zvibr = e−βω/2

∞ n=0

e−βωn =

e−βω/2 1 − e−βω

(4.95)

At 1000 K, for I2 , the term βω is 0.308 : the quantum regime is valid and this expression cannot be approximated. It has the value 3.23.

110

Appendix 4.2

d. Electronic Term

ze =

ge e−βεe

(4.96)

e

Owing to the large distances between electronic levels, one will only consider the fundamental level, four times degenerated for I, nondegenerated for I2 . At 1000 K, the term zeI is equal to 4 e−(Ed /2kB T ) = 4e−8.94 = 5.25 × 10−4 , whereas for the same choice of energy origin zeI2 = 1.

3. Prediction of the Action of Mass Constant Let us return to the deﬁnition relations (4.68) to (4.70) of Appendix 4.1, Kp = exp(−∆G0m /RT ) = exp(−

n

bi µ0i /kB T )

(4.97)

i=1

where the µ0i are the chemical potentials under the atmospheric pressure. For our example : Kp = exp[−(2µ0I − µ0I2 )/kB T ]

(4.98)

The chemical potentials are expressed versus the partition functions calculated above : Å 0 Å 0 ã ã 1 z I tr z I2 tr 0 0 (2µI − µI2 ) = −2[ln + ln zeI ] + ln + ln zrot + ln zvib kB T NI NI2 (4.99) Here the translation partition functions are calculated under the atmospheric pressure (standard pressure) ; the terms associated to the nuclear spins 2 ln(2J + 1) exactly compensate and ln ze I2 = 0. From the data of the previous section one obtains : z 0 I tr = 20.87 , NI z 0 I2 tr ln = 21.91 , NI2 ln

ln zeI = −7.56 ln zrot = 9.14 , ln zvibr = 1.17

Kp = e−5.6 3.7 × 10−3

Appendix 4.2

111

One can also predict, from the calculation of the temperature variation of Kp , a value for the enthalpy of the iodine dissociation reaction : ∆H d ln Kp = dT RT 2

(4.100)

i.e., d dT

Ç

−2µ0I + µ0I2 kB T

å =

∆H RT 2

(4.101)

∆H is thus deduced from the logarithmic derivatives versus T , at constant P , of the diﬀerent partition functions : the translation partition functions vary ã Å βω −1 . Consequently, like T 5/2 , zrot like T , zvibr like sinh 2 ã ò ã ò ï Å 0 ï Å 0 z I tr z I2 tr d d ∆H =2 + ln zeI − + ln zrot + ln zvibr ln ln RT 2 dT NI dT NI2 (4.102) ò ï ò ï 5 Ed 1 ω βω 5 − + + + (4.103) coth =2 2 2 2T 2kB T 2T T 2kB T 2 that is, Å ã 3RT βω ω ∆H = + N Ed − coth 2 2 2

(4.104)

The obtained value of ∆H is close to that of the dissociation energy per mole : ∆H = 152 kJ/mole. The Kp and ∆H values, calculated from spectroscopic data, are very close to those deduced from the experiment quoted in the Appendix 4.1 (Kp 3.4 · 10−3 at 1000 K, ∆H = 158 kJ/mole). The discrepancies come in particular from the approximations done in the decomposition of the degrees of freedom of the iodine molecule : we assumed that the vibration and rotation were independent (rigid rotator model) ; moreover, the potential well describing vibration is certainly not harmonic. Obviously, for an experiment performed at constant temperature T and volume Ω, the equilibrium constant Kv could have been calculated in a similar way and the free energy F would have been used. The variation of Kv with temperature provides the variation ∆U of internal energy in the chemical reaction.

4. Conclusion Through these examples, which were developed in detail, the contribution of Statistical Physics in the prediction of chemical reactions was shown. As just

112

Appendix 4.2

seen now, if the Statistical Physics concepts refer to very fundamental notions, this discipline (sometimes at the price of lengthy calculations !) brings useful data to solve practical problems.

Chapter 5

Indistinguishability, the Pauli Principle (Quantum Mechanics) The preceding chapter presented properties of the ideal gas which are valid in the limit where the average value, at temperature T , of the de Broglie wavelength is small with respect to the average distance between the particles in motion. Another way to express the same concept is to consider that the wave functions of two particles have a very small probability of overlapping in such conditions. From now on, we will work in the other limit, in which collisions between particles of the same nature are very likely. In the present chapter the Quantum Mechanics properties applicable to this limit will be recalled. In Quantum Mechanics courses (see for example the one by J.-L. Basdevant and J. Dalibard) it is stated that, for indistinguishable particles, the Pauli principle must be added to the general postulates of Quantum Mechanics. In fact here we are only re-visiting this principle but ideas are better understood when they are tackled several times ! In chapter 6 the consequences in Statistical Physics will be drawn on the properties obtained in the present chapter 5. In the introduction (§ 5.1) we will list very diﬀerent physical phenomena, associated with the property of indistinguishability, which will be interpreted at the end of the chapter in the framework of Statistical Physics. To introduce the issue, in § 5.2 some properties of the quantum states of two, then several, identical particles, will be recalled. In § 5.3, the Pauli principle and 113

114

Chapter 5. Indistinguishability, the Pauli Principle

its particular expression, the Pauli exclusion principle, will be stated ; the theorem of spin-statistics connection will allow one to recognize which type of statistics is applicable according to the nature of the considered particles and to distinguish between fermions and bosons. In § 5.4, the special case of two identical particles of spin 1/2 will be analyzed. These results will be extended to an arbitrary number of particles in § 5.5 and the description of a N indistinguishable particle state through the occupation numbers of its various energy levels will be introduced : this is much more convenient than using the N -particle wave function. It is this approach that will be used in the following chapters of this book. Finally in § 5.6 we will return to the examples of the introduction and will interpret them in the framework of the Pauli principle.

5.1

Introduction

It is possible to distinguish ﬁxed particles of the same nature, by specifying their coordinates : this is the case of magnetic moments localized on sites of a paramagnetic crystal, of hydrogen molecules adsorbed on a solid catalyst. They are then distinguishable through their coordinates. D

θ (1)

(2)

a)

b)

c)

Fig. 5.1: Two identical particles, a) represented by the wave packets which describe them in their center-of-mass frame, b) suﬀer a collision. c) After the collision, the wave function is spread in a ring-shaped region the radius of which increases with time. When a particle is detected in D, it is impossible to know whether it was associated with wave packet (1) or (2) prior to the collision. On the other hand, two electrons in motion in a metal, two protons in a nucleus, or more generally all particles of the same chemical nature, identical, in motion and thus likely to collide, cannot be distinguished by their physical properties : these identical particles are said to be indistinguishable. Indeed, as explained in Quantum Mechanics courses, if these particles, the evolution of which is described through Quantum Mechanics, do collide, their wave

Introduction

115

functions overlap (ﬁg. 5.1). Then, after the collision it is impossible to know whether (1 ) originated from (1) or from (2), since the notion of trajectory does not exist in Quantum Mechanics (Fig. 5.2). D

D (1 )

(2 )

(1)

(1) (2)

(2) (2 )

(1 ) a)

b)

Fig. 5.2: Representation in term of “trajectory” of the event described on Fig. 5.1. The two particles being identical, one cannot distinguish between a) and b). The quantum treatment of such systems refers to the Pauli principle. This principle allows the interpretation of a great variety of chemical and physical phenomena, among which are 1. the periodic classiﬁcation of the elements ; 2. the chemical binding ; 3. the ferromagnetism of some metals, like iron or cobalt, that is, their ability to evidence a macroscopic magnetization even in the absence of an external magnetic ﬁeld ; 4. the superconductivity of some solids, the electrical resistance of which strictly vanishes below a temperature called “critical temperature” ; 5. the special behavior of the isotope 42 He, which becomes superﬂuid for temperatures lower than a critical temperature Tc , equal to 2.17 K under the atmospheric pressure : this means that no viscosity is restraining its motion ; the isotope 32 He does not show such a property ; 6. the stimulated light emission and the laser eﬀect.

116

5.2

5.2.1

Chapter 5. Indistinguishability, the Pauli Principle

States of Two Indistinguishable Particles (Quantum Mechanics) General Case

When N particles are indistinguishable, the physics of the system and the N particle hamiltonian are unchanged by permutation of some of these particles. Consequently, the hamiltonian commutes with any permutation operator between these particles, so that they have a common basis of eigenvectors. It can be shown that any permutation of N particles can be decomposed into a product of exchanges of two particles. Thus it is helpful to begin by recalling the properties of the operator which exchanges the particles of a system consisting in particle 1 at r1 and particle 2 at r2 : Pˆ12 ψ(r1 , r2 ) = ψ(r2 , r1 )

(5.1)

The operator Pˆ12 must satisfy : (Pˆ12 )2 = 1

(5.2)

since after two exchanges one returns to the initial state. This implies that the eigenvalue eiα of Pˆ12 must fulﬁll the condition : e2iα = 1, i.e., eiα = ±1

(5.3)

eiα = +1, ψ(r2 , r1 ) = ψ(r1 , r2 )

(5.4)

i) If

the eigenstate is symmetrical in the exchange (or transposition) of particles 1 and 2. ii) If eiα = −1, ψ(r2 , r1 ) = −ψ(r1 , r2 )

(5.5)

the wave function is antisymmetrical in the transposition of the two particles.

5.2.2

Independent Particles

If the indistinguishable particles are independent, the total hamiltonian for all the particles is written as a sum of one-particle hamiltonians : ˆN = H

N i=1

ˆi h

(5.6)

States of two Indistinguishable Particles (Quantum Mechanics)

117

The eigenstates of each one-particle hamiltonian satisfy : ˆ i χn (ri ) = εn χn (ri ) (i = 1, 2) h

(5.7)

The index n corresponds to the considered quantum state, for example for the electron of a hydrogen atom n = 1s, 2s, 2p . . . ; the particle i is located at ri . The one-particle eigenstate at the lowest energy (fundamental state) satisﬁes ˆ 1 χ1 (r1 ) = ε1 χ1 (r1 ) h

(5.8)

The energy ε2 of the one-particle ﬁrst excited state is such that ˆ 1 χ2 (r1 ) = ε2 χ2 (r1 ) h

(5.9)

Since the particles are assumed to be independent, the eigenstates of the ˆ 12 can be easily expressed with respect to the onetwo-particle hamiltonian H particle solutions : ˆ 12 ψnn (r1 , r2 ) = εnn ψnn (r1 , r2 ) H

(5.10)

εnn = εn + εn and ψnn (r1 , r2 ) = χn (rn )χn (r2 )

(5.11)

with

Indeed remember that ˆ1 + h ˆ 2 )χn (r1 )χn (r2 ) = εn χn (r1 )χn (r2 ) + χn (r1 )εn χn (r2 ) (h = (εn + εn )χn (r1 )χn (r2 )

(5.12)

so that the two-particle energy is the sum of the one-particle energies and the two-particle wave function is the product of the one-particle wave functions. Any linear combination of wave functions corresponding to the same energy ˆ 12 . εn + εn is also a solution of H Here one is looking for eigenfunctions with a symmetry determined by the indistinguishability properties. In the two-particle fundamental state, each individual particle is in its fundamental state (ﬁg. 5.3), this two-particle state has the energy 2ε1 , it is nondegenerate. Its wave function satisﬁes ˆ 12 ψ11sym (r1 , r2 ) = 2ε1 ψ11sym (r1 , r2 ) H

(5.13)

It is symmetrical and can be written ψ11sym (r1 , r2 ) = χ1 (r1 )χ1 (r2 ) .

(5.14)

118

Chapter 5. Indistinguishability, the Pauli Principle

ε2

ε2

ε1

ε1

fundamental state

ﬁrst excited state

Fig. 5.3 : Occupation of two levels by two indistinguishable particles. In the two-particle ﬁrst excited state, the total energy is ε1 + ε2 . This state is doubly degenerate, as two distinct independent states can be built (particle 1 in ε1 , particle 2 in ε2 or the opposite). But since the particles are indistinguishable, one cannot tell which of the two particles is in the fundamental state. ˆ 12 and Pˆ12 . Thus one looks for solutions which are common eigenstates of H The acceptable solutions are : for the symmetrical ﬁrst excited state 1 ψ12sym (r1 , r2 ) = √ (χ1 (r1 )χ2 (r2 ) + χ2 (r1 )χ1 (r2 )) 2

(5.15)

for the antisymmetrical ﬁrst excited state 1 ψ12asym (r1 , r2 ) = √ (χ1 (r1 )χ2 (r2 ) − χ2 (r1 )χ1 (r2 )) 2

(5.16)

These wave functions (5.14) to (5.16) are determined, to a phase factor. Now consider the case of N particles. ˆ N , is then equal The total energy of the system, which is the eigenvalue of H to εN =

N

εni

(5.17)

i=1

The eigenfunction is a product of one-particle eigenfunctions, N

ψN (r1 , . . . , rN ) = Π χni (ri ) i=1

(5.18)

Any permutation of N particles can be decomposed into a product of exchanges of particles by pair, it thus admits the eigenvalues +1 or −1 : Pˆ ψ(r1 , r2 , . . . ri , . . . rN ) = σψ(r1 , r2 , . . . ri , . . . , rN )

(5.19)

Pauli Principle ; Spin-Statistics Connection

119

with σ = ±1. σ = +1 is associated with an “even” permutation σ = −1 is associated with an “odd” permutation Although the decomposition of the permutation into a product of exchanges is not unique, the value of σ is well determined. Among the eigenfunctions of the permutation operators, two types will be distinguished : either the wave function is completely symmetrical, i.e., it is unchanged by any permutation operator ; or it is a completely antisymmetrical eigenfunction which satisﬁes (5.19) and changes its sign for an odd permutation, but is not changed for an even permutation. The eigenfunctions, common ˆ N of the N indistinto the permutation operators and to the hamiltonian H guishable particles, will necessarily be of one of these two types.

5.3

5.3.1

Pauli Principle ; Spin-Statistics Connection Pauli Principle ; Pauli Exclusion Principle

The Pauli (1900-1958) principle is a postulate which is stated as follows : The wave function of a set of N indistinguishable particles obeys either of the two properties below : – either it is completely symmetrical, i.e., its sign does not change by permutation of two arbitrary particles : then these particles are called bosons ; – or it is completely antisymmetrical, and thus its sign is modiﬁed by permutation of two particles : these particles are fermions. The property veriﬁed by the wave function only depends on the type of the considered particles, it is independent of their number N and of the particular state Ψ they occupy. This principle is veriﬁed by all chemical types of particles. In particular, every time a new type of elementary particle was discovered, these new particles indeed satisﬁed the Pauli principle. A special case of the Pauli principle is the Pauli exclusion principle : Two fermions cannot be in the same quantum state. Indeed, if they were in the

120

Chapter 5. Indistinguishability, the Pauli Principle

same one-particle state Ψ, the antisymmetrical expression of the wave function would be : 1 ψ(1, 2) = √ [ψ(1)ψ(2) − ψ(2)ψ(1)] = 0 2

(5.20)

One now needs to determine which particles, called “fermions”, satisfy the Fermi-Dirac statistics, and which particles, the “bosons”, follow the BoseEinstein statistics. There is no intuitive justiﬁcation to the following theorem, which provides the answer to this question.

5.3.2

Theorem of Spin-Statistics Connection

The particles with an integral or zero spin (S/ = 0, 1, 2 . . . ) are bosons. The wave function for N bosons is symmetrical in the transposition of two of them. The particles with an half-integral spin (S/ = 1/2, 3/2, 5/2 . . . ) are fermions. The N -fermions wave function is antisymmetrical in the transposition of two of them. Among the elementary particles, electrons, protons, neutrons, neutrinos, muons have a spin equal to 1/2 ; the spin of photons is 1, the spin of mesons is 0 or 1. When these particles are combined into one or several nuclei, the total angular momentum takes into account all the orbital momenta and the spin intrinsic momenta. In Quantum Mechanics courses it is shown that the addition of two angular momenta J1 and J2 gives a parameter which has the properties of a new angular momentum, with a quantum number J varying between | J1 −J2 | and | J1 + J2 | by unity steps. If two integral or half-integral momenta are added, only integral values of J are obtained ; if an integral J1 is added to a half-integral J2 , the sum J is half-integral. It is the quantum number J which is identiﬁed with the spin of the ensemble. More generally if a composite particle contains an even number of elementary particles of half-integral spin, its total angular moment (or its total spin) is integral, this is a boson ; if it contains an odd number, its total angular moment is half-integral, and the composite particle is a fermion. For example, the atom 42 He has two electrons, two protons and two neutrons, i.e., six particles of half-integral spin : it is a boson. On the other hand, the atom 32 He is composed of two electrons, two protons but a single neutron, thus it is a fermion. The statistical properties of these two isotopes are very diﬀerent (see the Introduction of this chapter, example 5), whereas the che-

Case of Two Particles of Spin 1/2

121

mical properties, determined by the number of electrons that can participate in chemical bonds, are the same.

5.4

Case of Two Particles of Spin 1/2

Here the properties of the wave functions of two indistinguishable particles of spin 1/2 are recalled, this is in particular the case for two electrons. From the theorem of spin-statistics connection, this wave function must be antisymmetrical with respect to the transposition of these two particles. Now, in addition to their spin degree of freedom, these two particles also have an orbital degree of freedom concerning the space coordinates ; from Quantum Mechanics courses, the wave function for a single particle of this type can be decomposed into tensor products concerning the space and spin variables. The most general form of decomposition is : ψ(r1 , r2 ; σ1 , σ2 ) = ψ++ (r1 , r2 )| + + + ψ+− (r1 , r2 )| + − + ψ−+ (r1 , r2 )| − + + ψ−− (r1 , r2 )| − −

5.4.1

(5.21)

Triplet and Singlet Spin States

Let us ﬁrst assume, for simpliﬁcation, that this decomposition is reduced to a single term, i.e., ψ(r1 , r2 ; σ1 , σ2 ) = ψ(r1 , r2 ) ⊗ |σ1 , σ2 , with σ1 , σ2 = ±1/2

(5.22)

Since we are dealing with fermions, the two-particle wave function changes sign when particles 1 and 2 are exchanged. This sign change can occur in two diﬀerent ways : – either it is the space part ψ(r1 , r2 ) of the wave function which is antisymmetrical in this exchange, the spin function remaining unchanged ; – or it is |σ1 , σ2 which changes its sign in this exchange, the space function remaining unchanged. Consider the spin states. Each particle has two possible spin states, noted | ↑ or | ↓. There are thus four basis states for the set of two spins, which can be chosen as | ↑↑, | ↑↓, | ↓↑, | ↓↓ (the order of notation is : particle 1, then particle 2).

(5.23)

122

Chapter 5. Indistinguishability, the Pauli Principle

From these states one deduces four new independent states, classiﬁed according to their symmetry, that is, their behavior when exchanging the spins of the two particles. There are three symmetrical states, unchanged when the role of particles 1 and 2 is exchanged : ⎧ ⎪ | ↑↑ = |11 ⎪ ⎪ ⎪ ⎨ 1

√ | ↑↓ + | ↓↑ = |10 ⎪ ⎪ 2 ⎪ ⎪ ⎩ | ↓↓ = |1 − 1

(5.24)

They constitute the triplet state of total spin S = 1, with projections on the quantization axis Oz described by the respective quantum numbers Sz = +1, 0, −1 : the eigenvalue of the squared length of this spin is equal to S(S + 1)2 , i.e., 22 when the eigenvalue of the projection along Oz is Sz , i.e., , 0 or −. There remains an antisymmetrical state, which changes its sign in the exchange of the two spins, this is the singlet state 1

√ | ↑↓ − | ↓↑ = |00 2

(5.25)

in which S = Sz = 0. Even when the two-particle hamiltonian does not explicitly depend on spin, orbital wave functions with diﬀerent symmetries and possibly diﬀerent energy eigenvalues are associated with the triplet and singlet states : for example, in the special case treated in § 5.2.2 the fundamental state, with a symmetrical orbital wave function [see Eq. (5.14)], must be a spin singlet state, whereas in the ﬁrst excited state one can have a spin triplet. This energy splitting between the symmetrical and antisymmetrical orbital solutions is called the exchange energy or exchange interaction. The exchange interaction is responsible of ferromagnetism, as understood by Werner Heisenberg (1901-1976) as early as 1926 : indeed the dipolar interaction between two magnetic moments a fraction of a nm apart, each of the order of the Bohr magneton, is much too weak to allow the existence of an ordered phase of magnetization at room temperature. On the contrary, owing to the Pauli principle, in ferromagnetic systems the energy splittings between states with diﬀerent spins correspond to separations between diﬀerent orbital levels. These are of a fraction of an electron-volt, which is equivalent to temperatures of the order of 1000 K, since these states are the ones responsible for valence bindings.

Case of Two Particles of Spin 1/2

123

In a ferromagnetic solid, magnetism can originate from mobile electrons in metals like in Fe, Ni, Co, or from ions in insulating solids, as in the magnetite iron oxide Fe3 O4 , and the spin can diﬀer from 1/2. However this interaction between the spins, which relies on the Pauli principle, is responsible for the energy splitting between diﬀerent orbital states : one state in which the total magnetic moment is diﬀerent from zero, corresponding to the lower temperature state, and the other one in which it is zero.

5.4.2

General Properties of the Wave Function of Two Spin 1/2 Particles

It has just been shown that, because of the Pauli principle, space wave functions are associated with the triplet states, which are changed into their opposite in the exchange of the two particles (antisymmetrical wave function). To the singlet state is associated a symmetrical wave function. The most general wave function for two spin 1/2 particles will be written, for a decomposition onto the triplet and singlet states : A A A S |11 + ψ10 |10 + ψ1−1 |1 − 1 + ψ00 |00 ψ11 A The wave functions of the type ψ11 are spatial functions of r1 and r2 , the upper index recalls the symmetry of each component.

The same wave function can be expressed on the spin basis |σ1 , σ2 , accounting for expressions (5.24) and (5.25) of the triplet and singlet wave functions. One thus obtains : A ψ(r, r2 ; σ1 , σ2 ) = ψ11 | + + + A + ψ1−1 | − −

A ψA − ψS + ψS ψ10 √ 00 | + − + 10√ 00 | − + 2 2

(5.26)

It thus appears, by comparing to the general expression (5.21), that the Pauli principle dictates symmetry conditions on the spin 1/2 components : • ψ++ (r1 , r2 ) and ψ−− (r1 , r2 ) must be antisymmetrical in the exchange of particles 1 and 2 • ψ+− (r1 , r2 ) and ψ−+ (r1 , r2 ) do not have a speciﬁc symmetry, but their sum is antisymmetrical and their diﬀerence symmetrical.

124

5.5

Chapter 5. Indistinguishability, the Pauli Principle

Special Case of N Independent Particles ; Occupation Numbers of the States

In the general case of an N -identical-particle system, the Pauli principle dictates a form for the N -particle wave function and occupation conditions for the energy levels which are very diﬀerent in the case of fermions or bosons. In this part of the course, we are only interested by independent particles : this situation is simpler and already allows a large number of physical problems to be solved.

5.5.1

Wave Function

Fermions : We have seen that two fermions cannot be in the same quantum state and that their wave function must be antisymmetrical in the transposition of the two particles. We consider independent particles, the N -particle wave function of which takes the form of a determinant called “Slater determinant,” which indeed ensures the change of sign by transposition and the nullity if two particles are in the same state : χ1 (r1 ) χ2 (r1 ) χN (r1 ) .. ... ψ(r1 , . . . , rN ) ∝ ... . .. χ1 (rN ) . χN (rN )

(5.27)

If a state is already occupied by a fermion, another one cannot be added. A consequence is that at zero temperature all the particles cannot be in the minimum energy state : the diﬀerent accessible levels must be ﬁlled, beginning with the lowest one, each time putting one particle per level and thus each time moving to the next higher energy level. The process stops when the N th and last particle is placed in the last occupied level : in Chemistry this process is called the “Aufbau principle” or construction principle ; in solids, the last occupied state at zero temperature is called the Fermi level and is usually noted εF (see chapters 7 and 8). Bosons : On the other hand, the wave function for N bosons is symmetrical ; in particular it is possible to have all the particles on the same level of minimum energy at zero temperature.

Special Case of N Independent Particles ; Occupation Numbers

5.5.2

125

Occupation Numbers

The N -fermion (or N -boson) wave functions are complicated to write because, as soon as a particle is put in a state of given energy, this induces conditions on the available states for the other particles ; then the antisymmetrization of the total wave function has to be realized for the fermions (or the symmetrization for the bosons). Since these particles are indistinguishable, the only interesting datum in Statistical Physics is the number of particles in a deﬁned quantum state. Instead of writing the N -particle wave function, which will be of no use in what follows, we will described the system as a whole by the occupation numbers of its diﬀerent states, this is the so-called “Fock basis” : these one-particle states correspond to distinct εk , which can have the same energy but diﬀerent indexes if the state is degenerate (for example two states of the same energy and diﬀerent spins will have diﬀerent indexes k and k ) ; the states are ordered by increasing energies (Fig. 5.4). εk

nk

ε3

n3

ε2

n2

ε1

n1

Fig. 5.4 : Deﬁnition of the occupation numbers of the one-particle states. The notation |n1 ,2 , . . . nk , . . . will express the N -particle state in which n1 particles occupy the state of energy ε1 , n2 that of energy ε2 , . . . nk the state of energy εk . One must satisfy n1 + n2 + . . . + nk + . . . = N

(5.28)

the total number of particles in the system. Under these conditions, the total energy of the N -particle system is given by n1 ε1 + n2 ε2 + . . . + nk εk + . . . = EN

(5.29)

From the Pauli principle • for fermions : nk = 0 or 1

(5.30)

• for bosons

(5.31)

: nk = 0, 1, 2 . . . ∞

126

Chapter 5. Indistinguishability, the Pauli Principle

This description will be used in the remainder of the course. The whole physical information is included in the Fock notation. Indeed one cannot tell which particle is in which state, this would be meaningless since the particles are indistinguishable. The only thing that can be speciﬁed is how many particles are in each one-particle state.

5.6

Return to the Introduction Examples

Now we are able to propose a preliminary interpretation of the physical phenomena described in the introduction, § 5.1, some of which will be analyzed in more detail in the following chapters :

5.6.1

Fermions Properties

1. In the periodic classiﬁcation of the elements, the atomic levels are ﬁlled with electrons which are fermions. In a given orbital level only two electrons of distinct spin states can be placed, i.e., according to the Pauli principle they are in the singlet spin state. 2. When a chemical bond is created, the electrons coming from the two atoms of the bond go into the atomic levels originating from the coupling between the levels of the separated atoms. The resulting bonding and antibonding states must also be populated by electrons in diﬀerent spin states, thus in the singlet spin state. 3. The ferromagnetism arises, as already mentioned, from the energy difference between the singlet and the triplet spin states for a pair of electrons. If the fundamental state presents a macroscopic magnetization, it must be made from microscopic states with a nonzero spin. In fact, in a ferromagnetic metal such as iron or nickel, the spins of two iron nuclei couple through conduction electrons which travel from nucleus to nucleus.

5.6.2

Bosons Properties

4. In superconductivity the electrons of a solid couple by pairs due to their interaction with the solid vibrations. These pairs, called “Cooper pairs,” are bosons which gather into the same fundamental state at low temperature, thus producing the phenomenon of superconductivity. 5. In the same way the 4 He atoms are bosons, which can occupy the same quantum state below a critical temperature (see chapter 9). Their ma-

Return to the Introduction Examples

127

croscopic behavior expresses this collective property, whence the absence of viscosity. 6. Photons are also bosons and the stimulated emission, a speciﬁc property of bosons in nonconserved number, will be studied in § 9.2.4.

Summary of Chapter 5 In Quantum Mechanics, the hamiltonian describing the states of identical particles is invariant under a permutation of these particles, the states before and after permutation are physically indistinguishable. This means that the hamiltonian commutes with any permutation operator : one ﬁrst looks for the permutations eigenstates, which are also eigenstates of the N -particle hamiltonian. In the case of two indistinguishable particles the symmetrical and antisymmetrical eigenstates have been described. These states have a simple expression in the case of independent particles. Any permutation of N particles can be decomposed into a product of exchanges (or transpositions) of particles by pair : a permutation is thus either even or odd. The Pauli principle postulates that the wave functions for N identical particles can only be, according to the particle’s nature, either completely symmetrical (bosons) or completely antisymmetrical (fermions) in a permutation of the particles. In particular, because their wave function is antisymmetrical, two fermions cannot occupy the same quantum state (“Pauli exclusion principle”). The theorem of spin-statistics connection speciﬁes that the particles of halfintegral total spin, the fermions, have an antisymmetrical wave function ; the particles of integral or zero total spin, the bosons, have a completely symmetrical wave function. In the case of two spin-1/2 particles, the triplet spin state, symmetrical, and the singlet spin state, antisymmetrical, have been described. The spin component of the wave function has to be combined with the orbital part, the global wave function being antisymmetrical to satisfy the Pauli principle. In Statistical Physics, to account for the Pauli principle, the state of N particles will be described using the occupation numbers of the various energy 129

130

Summary of Chapter 5

levels, rather than by its wave function. A given state can be occupied at most by one fermion, whereas the number of bosons that can be located in a given state is arbitrary.

Chapter 6

General Properties of the Quantum Statistics In chapter 5 we learnt the consequences, in Quantum Mechanics, of the Pauli principle which applies to indistinguishable particles : they will now always be taken as independent, that is, without any mutual interaction. We saw that it is very diﬃcult to express the conditions stated by the Pauli principle on the N -particle wave function ; on the other hand, a description through occupation numbers |n1 , n2 , . . . , nk , . . . of the diﬀerent quantum states ε1 , ε2 , . . . εk . . . accessible to these indistinguishable particles is much more convenient. In fact we assume that the states are nondegenerate and if there are several states with the same energy, we label them with diﬀerent indexes. Besides, it was stated that there exists two types of particles, the fermions with a half-integral spin, which cannot be more than one in any given one-particle state εk , whereas the bosons, with an integral or zero spin, can be in arbitrary number in any one single-particle quantum state. In the present chapter we return to Statistical Physics (except in § 6.4, which is on Quantum Mechanics) : after having presented the study technique applicable to systems referring to Quantum Statistics, we determine the average number of particles, at a given temperature, on an energy state of the considered system : this number will be very diﬀerent for fermions and bosons. This chapter presents general methods and techniques to analyze, using Statistical Physics, the properties of systems of indistinguishable particles following the Quantum Statistics of either Fermi-Dirac or Bose-Einstein, while the following chapters will study these statistics more in detail on physical examples.

131

132

Chapter 6. General Properties of the Quantum Statistics

In § 6.1, we will show how the statistical properties of indistinguishable independent particles are more conveniently described in the grand canonical ensemble. In § 6.2 we will explain how the grand canonical partition function ZG is factorized into terms, each one corresponding to a one-particle quantum state. In § 6.3 the Fermi-Dirac and the Bose-Einstein distributions will be deduced, which give the average number of particles of the system occupying a quantum state of given energy at temperature T , and the associated thermodynamical potential, the grand potential, will be expressed. The § 6.4 will be a (long) Quantum Mechanics parenthesis, it will be shown there that a macroscopic system, with extremely close energy levels, is conveniently described by a density of states, that will be calculated for a free particle. In § 6.5, this density of states will be used to obtain the average value, at ﬁxed temperature, of physical parameters in the case of indistinguishable particles. Finally, in § 6.6, the criterion used in § 4.3 to deﬁne the domain of Classical Statistics will be justiﬁed and it will be shown that the Maxwell-Boltzmann Classical Statistics is the common limit of both Quantum Statistics when the density decreases or the temperature increases.

6.1

Use of the Grand Canonical Ensemble

In chapter 5 we learnt the properties requested for the wave function of N indistinguishable particles, when these particles are fermions or when they are bosons. (From the present chapter until the end of the book, the study will be limited to indistinguishable and independent particles.) Now we look for the statistical description of such a system in thermal equilibrium at temperature T and ﬁrst show, on the example of two particles, that the use of the canonical ensemble is not convenient.

6.1.1

Two Indistinguishable Particles in Thermal Equilibrium : Statistical Description in the Canonical Ensemble

Consider a system of two indistinguishable particles. The one-particle energy levels are noted ε1 , . . . , εk , . . . , and are assumed to be nondegenerate (a degenerate level is described by several states at the same energy but with diﬀerent indexes). The canonical partition function for a single particle at temperature T is written : 1 (6.1) e−βεk with β = Z1 = kB T k

Use of the Grand Canonical Ensemble

133

Now consider the two particles. For a given microscopic conﬁguration, the total energy is εi + εj . After having recalled the results for two distinguishable particles, we will have to separately consider the case of fermions and that of bosons. We have seen in § 2.4.5 that for distinguishable particles Z2disc = (Z1 )2

(6.2)

since there is no limitation to the levels occupation by the second particle. Fermions cannot be more than one to occupy the same state. Consequently, as soon as the level of energy εi is occupied by the ﬁrst particle, the second one cannot be on it. Thus 1 −β(εi +εj ) −β(εi +εj ) Z2fermions = e = e (6.3) 2 i<j i=j

Let us compare Z2fermions to (Z1 )2 : 2 −βεi −βεj (Z1 ) = e e e−β2εi + 2 e−β(εi +εj ) = i

(Z1 )2 =

j

i

(6.4)

i<j

e−β2εi + 2Z2fermions

(6.5)

i

In the case of bosons, in addition to the conﬁgurations possible for the fermions, the two particles can also lie on the same level : e−β(εi +εj ) (6.6) Z2bosons = i≤j 2

Let us relate (Z1 ) to Z2bosons : (Z1 )2 = Z2bosons +

1 −β(εi +εj ) e 2

(6.7)

i=j

(Z1 )2 = Z2bosons + Z2fermions

(6.8)

It was easy to determine the levels available for the particles in this example because they were only two. For three particles one must then consider which restrictions appear for the third one because of the levels occupied by the ﬁrst two particles. For N particles one should again proceed step by step, but this method is no longer realistic !

6.1.2

Description in the Grand Canonical Ensemble

To avoid complicated conditions on the occupation numbers nk one will work in the grand canonical ensemble, assuming in a ﬁrst step that the total number

134

Chapter 6. General Properties of the Quantum Statistics

of particles of the system is arbitrary. Thus one will calculate ZG and the grand potential A ; then the value of the Lagrange parameter α = βµ (or that of the chemical potential µ) will be ﬁxed in such a way that the average number of particles N , deduced from the statistical calculation, coincides with the real number of particles in the system. We know (§ 2.5.4) that the √ relative ﬂuctuation on N introduced by this procedure is of the order of 1/ N , that is, of the order of a few 10−12 for N of the order of the Avogadro number.

6.2

6.2.1

Factorization of the Grand Partition Function Fermions and Bosons

By deﬁnition ∞

ZG (α, β) =

∞

eαN ZN (β) =

N =0

eβµN ZN (β)

(6.9)

N =0

where N is one of the numbers of particles and where α and the chemical potential are related by α = βµ. The canonical partition function ZN (β) corresponding to N particles is given by

ZN (β) =

e−βEn

(6.10)

n

Let us introduce the occupation numbers nk of the quantum states εk :

ZG =

∞

βµ

e

Ñ

nk

k

N =0

−β

e

é

nk εk

k

(6.11)

nk =N

with

nk = N

(6.12)

k

and the energy for this conﬁguration of the N particles being given by k

nk εk = EN

(6.13)

Factorization of the Grand Partition Function

135

As the sum is performed over all the total numbers N and all the particle distributions in the system, ZG is also written βµ nk −β nk εk k ZG = e k (6.14) {|n1 ...nk >}

=

β

e

nk (µ−εk )

k

(6.15)

{|n1 ...nk >}

In the latter expression, one separates the contribution of the state ε1 : this is the term eβ(µ−ε1 )n1 (6.16) n1

which multiplies the sum concerning all the other states, with n1 taking all possible values. One then proceeds state after state, so that ZG is written as a product of factors : ∞ β(µ−εk )nk ZG = e (6.17) k=1

nk

Each factor concerns a single state εk and its value depends on the number of particles which can be located in this state, i.e., of the nature, fermions or bosons, of the considered particles. Note that, since ZG is a product of factors, ln ZG is a sum of terms, each concerning a one-particle state of the type εk .

6.2.2

Fermions

In this case, the quantum state εk is occupied by 0 or 1 particle : there are only two terms in nk , the contribution of the state εk in ZG is the factor 1 + exp β(µ − εk ), whence ZG fermions =

∞

(1 + eβ(µ−εk ) )

(6.18)

k=1

6.2.3

Bosons

The state εk can now be occupied by 0, 1, 2,... particles, so that nk varies from zero to inﬁnity. The contribution in ZG of the state εk is thus the factor ∞

eβ(µ−εk )nk =

nk =0

1 1 − eβ(µ−εk )

(6.19)

136

Chapter 6. General Properties of the Quantum Statistics

This geometrical series can be summed only if exp(α − βεk ) is smaller than unity, a condition to be fulﬁlled by all the states εk . One will thus have to verify that µ < ε1 , where ε1 is the one-particle quantum state with the lowest energy, i.e., the fundamental state. This condition is also expressed through α = βµ as α < βε1 . Then the grand partition function takes the form : ZG bosons =

∞ k=1

6.2.4

1 1−

eβ(µ−εk )

(6.20)

Chemical Potential and Number of Particles

The grand canonical partition function ZG is associated with the grand potential A such that A = −kB T ln ZG

(6.21)

As speciﬁed in § 3.5.3, the partial derivatives of the grand partition function provide the entropy, the pressure and the average number of particles N : dA = −SdT − P dΩ − N dµ

(6.22)

In particular the constraint on the total number of particles of the system is expressed from : Å ã ∂A N =− (6.23) ∂µ T,Ω

6.3

Average Occupation Number ; Grand Potential

From the grand partition function ZG one deduces average values at temperature T of physical parameters of the whole system, but also of the state εk . In particular the average value of the occupation number nk of the state of energy εk is obtained through nk =

nk eβ(µ−εk )nk 1 ∂ ln ZG β(µ−ε )n =− k k β ∂εk e

nk

nk

(6.24)

Average Occupation Number ; Grand Potential

137

For fermions, the average occupation number, the so-called “Fermi-Dirac distribution,” is given by nk F D =

1 1 = βε −α e k +1 eβ(εk −µ) + 1

with α = βµ

(6.25)

For bosons, this average number is “the Bose-Einstein” distribution which takes the form nk BE =

1 1 = βε −α e k −1 eβ(εk −µ) − 1

(6.26)

Note : Remember the sign diﬀerence in the denominator between fermions and bosons ! Its essential physical consequences are developed in the next chapters. The average occupation numbers are related to factors in the grand partition function ZG : indeed one will verify that 1 = 1 + eβ(µ−εk ) 1 − nk F D 1 1 + nk BE = 1 − eβ(µ−εk )

(6.27) (6.28)

which are the respective contributions of the level εk to ZG , in the cases of fermions or bosons. These occupation numbers yield the average values, at given temperature, thus at ﬁxed β, of physical parameters. Thus the average number of particles is related to α and β, which appear in f (ε), through N =

nk

(6.29)

k

The total energy of the system at temperature T is obtained from U=

εk nk

(6.30)

k

The grand potential A is expressed, like the ZG factors, versus the occupation numbers of the states of energy εk : A = +kB T

ln(1 − nk ) fermions

(6.31)

ln(1 + nk ) bosons

(6.32)

k

A = −kB T

k

138

6.4

Chapter 6. General Properties of the Quantum Statistics

Free Particle in a Box ; Density of States (Quantum Mechanics)

The diﬀerent physical parameters, like the average number of particles N , the internal energy U , the grand potential A, were expressed above as sums of contributions arising from the discrete one-particle levels. Now we will ﬁrst recall what these states are for a free particle conﬁned within a volume of macroscopic characteristic dimensions, i.e., of the order of a mm or a cm (§ 6.4.1) : such dimensions are extremely large with respect to atomic distances, so that the characteristic energy splittings are very small (§ 6.4.1a). The boundary conditions (§ 6.4.1b) of the wave function imply quantization conditions, that we are going to express in two diﬀerent ways. However, for both quantization conditions the same energy density of states D(ε) is deﬁned (§ 6.4.2), such that the number of allowed states between energies ε and ε + dε is equal to D(ε) dε.

6.4.1

Quantum States of a Free Particle in a Box

a) Eigenstates By deﬁnition a free particle feels no potential energy (except the one expressing the possible conﬁnement) We ﬁrst consider that the free particle can move through the entire space. Its hamiltonian is given by pˆ 2 ˆ h= 2m

(6.33)

We are looking for its stationary states (see a course on Quantum Mechanics) : the eigenstate |ψ and the corresponding energy ε satisfy pˆ 2 |ψ = ε|ψ 2m

(6.34)

i.e., 2 ∆ψ(r) = εψ(r) 2m The time-dependent wave function is then deduced : −

Ψ(r, t) = ψ(r) exp(−iεt/)

(6.35)

(6.36)

It is known that the three space variables separate in eq. (6.35) and that it is enough to solve the one-dimension Schroedinger equation −

2 d2 ψ(x) = εx ψ(x) 2m dx2

(6.37)

Free Particle in a Box ; Density of States

139

which admits two equivalent types of solutions : i) either 1 ψ(x) = √ exp ikx x Lx with kx > 0, < 0 or zero. The time-dependent wave function Å ã 1 εx t Ψ(x, t) = √ exp i kx x − Lx

(6.38)

(6.39)

2 kx2 = ωx is a progressive wave : when the time t increases, with εx = 2m a state of given phase propagates toward increasing x’s for kx > 0, toward decreasing x’s for kx < 0 ; the space wave function is a constant for kx = 0. ii) or ψ(x) = A cos kx x + B sin kx x

(6.40)

where A and B are constants. The time-dependent wave function ã Å εx t Ψ(x, t) = (A sin kx x + B cos kx x) exp −i

(6.41)

corresponds to the same energy εx or the same pulsation ωx as in (6.39). The solution of progressive-wave type (6.39) is found again if one chooses √ A = 1/ Lx , B = iA. For real coeﬃcients A and B, when changing the space origin in x there is indeed separation of the space and time variables in (6.41). Then a state of given space phase does not propagate in time, such a wave is stationary. At three dimensions the kinetic energy terms along the three coordinates add up in the hamiltonian ; the energy is equal to ε = εx + εy + εz

(6.42)

and the wave function is the product of three wave functions, each concerning a single coordinate. Now we are going to express, through boundary conditions on the wave function, that the particle is conﬁned within the volume Ω. This will yield quantization conditions on the wave vector and the energy.

140

Chapter 6. General Properties of the Quantum Statistics

b) Boundary Conditions The volume Ω containing the particles has a priori an arbitrary shape. However, it can be understood that each particle in motion is most of the time at a large distance from the container walls, so that the volume properties are not sensitive to the surface properties as soon as the volume is large enough. It is shown, and we will admit it, that the Statistical Physics properties of a macroscopic system are indeed independent of the shape of the volume Ω. For convenience we will now assume that Ω is a box (Lx , Ly , Lz ), with macroscopic dimensions of the order of a mm or a cm. We consider a one-dimension problem, on a segment of length Lx . The presence probability of the particle is zero outside the interval [O, Lx ], in the region where the particle cannot be found. There are two ways to express this condition : i) Stationary boundary conditions : To indicate that the particle cannot leave the interval, one assumes that in x = 0 and x = Lx potential barriers are present, which are inﬁnitely high and thus impossible to overcome. Thus the wave function vanishes outside the interval [O, Lx ] and also at the extremities of the segment, to ensure its continuity at these points ; since it vanishes at x = 0, it must be of the form ψ(x) = A sin kx x (Fig. 6.1). The cancellation at x = Lx yields kx Lx = nx π

(6.43)

where nx is an integer, so that π Lx π The wave vector is an integer multiple of , it is quantized. Lx kx = nx

For such values of kx , the time-dependent wave function is given by ã Å εx t Ψ(x, t) = A sin kx x exp −i

(6.44)

(6.45)

One understands that the physically distinct allowed values are restricted to nx > 0 : indeed taking nx < 0 is just changing the sign of the wave function, which does not change the physics of the problem. The value nx = 0 is to be rejected since a wave function has to be normalized to unity and thus cannot cancel everywhere. In the same way ky and kz are quantized by the conditions of cancellation of the wave function on the surface : ﬁnally the allowed vectors k for the particle conﬁned within the volume Ω are of the form ã Å k = nx π , ny π , nz π (6.46) Lx Ly Lz

Free Particle in a Box ; Density of States

141

ψ(x)

x O

Lx

Fig. 6.1: The free-particle wave function is zero outside the interval ]O, Lx [. the three integers nx , ny , nz being strictly positive. In the space of the wave vectors (Fig. 6.2), the extremities ã are on a rectangular lattice Å of these vectors π π π k . The unit cell built on these deﬁned by the three vectors i, j, Lx Ly Lz π3 π3 . = vectors is Lx Ly Lz Ω kz

L

x

π

π Ly π Lz

ky

kx Fig. 6.2: Quantization of the k-space for boundary conditions of the stationary-wave type. Only the trihedral for which the three k-components are positive is to be considered. ii) Periodical boundary [Born-Von Kármán (B-VK)] conditions : Here again one considers that when a solid is macroscopic, all that deals with its surface has little eﬀect on its macroscopic physical parameters like its pressure, its temperature. The solid will now be closed on itself (Fig. 6.3) and this will not noticeably modify its properties : one thus suppresses surface

142

Chapter 6. General Properties of the Quantum Statistics

eﬀects, of little importance for a large system. Then at a given x the properties are the same, and the wave function takes the same value, whether the particle arrives at this point directly or after having traveled one turn around before reaching x : ψ(x + Lx ) = ψ(x)

(6.47)

Lx O

x

Fig. 6.3 : The segment [0, Lx ] is closed on itself. This condition is expressed on the progressive-wave type wave function (6.39) through exp ikx (x + Lx ) = exp ikx x

(6.48)

for any x, so that the condition to be fulﬁlled is kx Lx = nx 2π 2π that is, kx = nx Lx

(6.49) (6.50)

where nx is integer. The same argument is repeated on the y and z coordinates. The wave vector k is again quantized, with basis vectors twice larger than for the condition i) above for stationary waves, the allowed values now being ã Å k = n 2π n 2π n 2π (6.51) x Lx y Ly z Lz The integers nx , ny , nz are now positive, negative or null (Fig. 6.4). Indeed the time-dependent wave function is written, to a phase, 2 k 2 1 Ψ(r, t) = √ exp i(k.r − ωt), with ω = ε = 2m Ω

(6.52)

Changing the sign of k, thus of the integers ni , produces a new wave which propagates in the opposite direction and is thus physically diﬀerent. If one of these integers is zero, this means that the wave is constant along this axis.

Free Particle in a Box ; Density of States

143

kz 2π Ly

L

x

2π

2π Lz

ky

kx Fig. 6.4: Quantization of the k-space for the periodic (B-VK) boundary conditions. Note that each dimension of the elementary box is twice that corresponding to the stationary boundary conditions (Fig. 6.2), and that now the entire space should be considered.

6.4.2

Density of States

Under these quantization conditions what is the order of magnitude of the obtained energies, and of the spacing of the allowed levels. As an example consider the framework of the periodic boundary conditions. The quantized allowed energies are given by ñ Å ã2 Å ã2 Å ã2 ô 2 2 2π 2 2π 2 2π (6.53) + ny + nz n ε= 2m x Lx Ly Lz For Lx for the order of 1 mm, a single coordinate and a mass equal to that of the proton, 2 2m

Å

2π Lx

ã2 = 13.6 eV.

ã2 Å (2π)2 0.5 × 10−10 ∼ . = 8 × 10−16 eV 1840 10−3

(6.54)

For a mass equal to that of the electron, the obtained value, 1840 times larger, is still extremely small at the electron-volt scale ! The spacing between the considered energy levels is thus very small as compared to any realistic experimental accuracy, as soon as Lx is macroscopic, that is, large with respect to atomic distances (the Bohr orbit of the hydrogen atom is associated with energies in the electron-volt range, see the above es-

144

Chapter 6. General Properties of the Quantum Statistics

timation). If the stationary boundary conditions had been chosen, the energy splitting would have been four times smaller. The practical question which is raised is : “What is the number of available quantum states in a wave vector interval d3k around k ﬁxed, or in an energy interval dε around a given ε ?” It is to answer this question that we are going to now deﬁne densities of states, which summarize the Quantum Mechanics properties of the system. It is only in a second step (§ 6.5) that Statistical Physics will come into play through the occupation factors of these accessible states. a) Wave Vector and Momentum Densities of States (B-VK Conditions) : One is estimating the number of quantum states dn of a free particle, conﬁned within a volume Ω, of wave vector between k and k + d3k, the volume d3k (2π)3 . The wave vector density of being assumed to be large as compared to Ω states Dk (k) is deﬁned by dn = Dk (k)d3k

(6.55)

We know that the allowed states are uniformly spread in the k-space and that, in the framework of the B-VK periodic boundary conditions, for each (2π)3 there is a new quantum state (Fig. 6.4) : indeed a elementary volume Ω parallelepiped has eight summits, shared between eight parallelepipeds. The searched number dn is thus equal to dn =

d3k Ω 3 = d k, (2π)3 /Ω (2π)3

(6.56)

which determines Dk (k) =

Ω (2π)3

(6.57)

The above estimation has been done from the Schroedinger equation solutions related to the space variable r (orbital solutions). If the particle carries a spin s, to a given solution correspond 2s + 1 states with diﬀerent spins, degenerate in the absence of applied magnetic ﬁeld, which multiplies the above value of the density of states [Eq. (6.57)] by (2s + 1). Thus for an electron, of spin 1/2, there are two distinct spin values for a given orbital state. For a free particle, its wave vector and momentum are related by p = k d3 p = 3 d3k

(6.58)

Free Particle in a Box ; Density of States

145

The number dn of states for a free particle of momentum vector between p and p + d3 p, conﬁned within the volume Ω, is written, from eq. (6.56), dn =

p Ω Ω d3 = 3 d3 p = Dp ( p)d3 p 3 3 (2π) h

(6.59)

Here the (three-dimensional) momentum density of states has been deﬁned through Dp ( p) =

Ω h3

(6.60)

Note that the distance occupied by a quantum state, on the momentum axis of the phase space corresponding to the one-dimension motion of a single h . This is consistent with the fact that a cell of this phase space particle, is Lx has the area h and that here ∆x = Lx , since, the wave function being of plane-wave type, it is delocalized on the whole accessible distance. b) Three-Dimensional Energy Density of States (B-VK Conditions) : The energy of a free particle only contains the kinetic term inside the allowed volume, it only depends of its momentum or wave vector modulus : ε=

2 k 2 p2 = 2m 2m

(6.61)

All the states corresponding to a given … energy ε are, in the k-space, at the 2mε surface of a sphere of radius k = . If one does not account for the 2 particle spin, the number of allowed states between the energies ε and ε + dε is the number of allowed values of k between the spheres of radii k and k + dk, 2 kdk. (For clarity in the sketch, Fig. dk being related to dε through dε = m 6.5 is drawn in the case of a two-dimensional problem.) The concerned volume is 4πk 2 dk, the number dn of the searched states, without considering the spin variable, is equal to Ω Ω 4πk 2 dk = 4πk.kdk (2π)3 (2π)3 … Ω 2mε m = 2 dε 2π 2 2

(6.62) (6.63)

Thus a density of states in energy is obtained, deﬁned by dn(ε) = D(ε) dε, with Å ã √ (2s + 1) 2m 3/2 (6.64) D(ε) = CΩ ε with C = 4π 2 2

146

Chapter 6. General Properties of the Quantum Statistics

ky ε+ ε

dε

kx O

2π Lx

2π Ly Fig. 6.5: Number of states of energy between ε and ε + dε, (two-dimensional problem). This expression of the energy density of states, which includes the spin degeneracy, is valid for any positive energy ; the density of states vanishes for ε < 0, since then no state can exist, the energy being only of kinetic origin for the considered free particle. c) Densities of States (Stationary-Waves Conditions) We have seen that, in the case of stationary-wave conditions, all the physically distinct solutions are obtained when restricting the wave vectors choice to those with all three positive components. The volume corresponding to the states of energies between ε and ε + dε must thus be limited to the corresponding trihedral of the k-space. This is the 1/8th of the space between the two 1 spheres of radii k and k + dk, i.e., 4πk 2 dk, dk being still related to dε by 8 dε =

2 kdk m

(6.65)

However the wave vector density of states is eight times larger for the stationary-wave quantization condition than for the B-VK condition (the allowed points of the k-space are twice closer in the x, y and z directions). Then one obtains a number of orbital states, without including the spin degeneracy, of dn =

Ω1 4πk 2 dk π3 8

(6.66)

Fermi-Dirac Distribution ; Bose-Einstein Distribution

147

The density D(ε) thus has exactly the same value than in b) when the spin degeneracy (2s + 1) is introduced. One sees that both types of quantization condition are equivalent for a macroscopic system. In most of the remaining parts of this course, the B-VK conditions will be used, as they are simpler because they consider the whole wave vector space. Obviously, in speciﬁc physical conditions, one may have to use the stationary-wave quantization conditions. Notes : In presence of a magnetic potential energy, which does not aﬀect the orbital part of the wave function, the density of states Dk (k) is unchanged. On the contrary, D(ε) is shifted (see § 7.1.3). As an exercise, we now calculate the energy density of states for free particles in the two- or one-dimension space. The relations between ε and k, between dε and dk are not modiﬁed [(6.61) and (6.65)]. – The two-dimension elementary area is 2πkdk, the k -density of states Lx Ly is given by without taking the spin in account and the energy (2π)2 density of states is a constant. – In one dimension, the elementary length is 2dkx (kx and −kx correspond L to the same energy, whence the factor 2), the kx -density of states is 2π 1 without spin, and D(ε) varies like √ . D(ε) diverges for ε tending to ε zero but its integral, which gives the total number of accessible states from ε = 0 to ε = ε0 , is ﬁnite.

6.5

Fermi-Dirac Distribution ; Bose-Einstein Distribution

We have just seen that, for a macroscopic system, the accessible energy levels are extremely close and have deﬁned a density of states, which is a discrete function of energy ε : in the large volume limit, it is as if D(ε) was a function of the continuous variable ε. Let us now return to Statistical Physics in the framework of the large volumes limit. From the average number of occupation nk (see § 6.3) of an energy level εk , one deﬁnes the energy function “ occupation factor” or “distribution,” of Fermi-Dirac or of Bose-Einstein according to the nature of the considered

148

Chapter 6. General Properties of the Quantum Statistics

particles : 1

fF D (ε) =

eβ(ε−µ)

+1

=

1 eβε−α

+1

1 1 = eβ(ε−µ) − 1 eβε−α − 1

fBE (ε) =

(6.67) (6.68)

Recall that α and the chemical potential µ are related by α = βµ with 1 β= . kB T The consequences of the expression of their distribution will be studied in details in chapter 7 for free fermions and in chapter 9 for bosons.

6.6

Average Values of Physical Parameters at T in the Large Volumes Limit

On the one hand, the energy density of states D(ε) summarizes the Quantum Mechanical properties of the system, i.e., it expresses the solutions of the Schroedinger equation for an individual particle. On the other hand, the distribution f (ε) expresses the statistical properties of the considered particles. To calculate the average value of a given physical parameter, one writes that in the energy interval between ε and ε + dε, the number of accessible states is D(ε)dε, and that at temperature T and for a parameter α or a chemical potential µ, these levels are occupied according to the distribution f (ε) (of Fermi-Dirac or Bose-Einstein). Thus +∞ D(ε)f (ε) dε (6.69) N= −∞ +∞

U=

εD(ε)f (ε) dε

(6.70)

−∞

The temperature and the chemical potential are included inside the distribution expression. In fact the ﬁrst equation allows one to determine the chemical potential value. To obtain the grand potential, one generalizes to a large system relations (6.31) and (6.32) that were established for discrete parameters : +∞ A = +kB T D(ε) ln(1 − f (ε)) dε for fermions (6.71) −∞ +∞

A = −kB T

D(ε) ln(1 + f (ε)) dε for bosons −∞

(6.72)

Common Limit of the Quantum Statistics

6.7 6.7.1

149

Common Limit of the Quantum Statistics Chemical Potential of the Ideal Gas

In the case of free particles (fermions or bosons) of spin s, conﬁned in a volume Ω, one expresses the number of particles (6.69) using the expression (6.64) of the density of states D(ε) : √ εdε (2s + 1)Ω(2m)3/2 +∞ (6.73) N= 2 3 βε−α 4π e ∓1 0 where, in the denominator, the sign − is for bosons and the sign + for fermions. If the temperature is ﬁnite, the dimensionless variable of the integral is βε = x, so that this relation can be written : √ ã Å N 2mkB T 3/2 (2s + 1) +∞ xdx = (6.74) x−α ∓ 1 Ω 2 4π 2 e 0 The physical parameters are then put into the left member, which depends of the system density and temperature as √ Å ã3/2 h2 1 1 N 4π 2 N π 1 N (λth )3 = constant λ3th = Ω (2π)3 2s + 1 23/2 mkB T Ω 2 2s + 1 Ω (6.75) The thermal de Broglie wavelength λth appears, as deﬁned in § 4.3. The quantity (6.75) must be equal to the right member integral, which is a function of α such that ∞ √ xdx (6.76) u∓ (α) = x−α e ∓1 0 √ In fact (6.75) expresses, to the factor π/2(2s+1), the cube of the ratio of the thermal de Broglie wavelength λth to the average distance between particles d = (Ω/N )1/3 . It is this comparison which, in § 4.3, had allowed us to deﬁne the limit of the classical ideal gas model. Since (6.75) and (6.76) are equal, one obtains : √ 1 πN (λth )3 (6.77) u∓ (α) = 2s + 1 2 Ω The integral u∓ (α) can be numerically calculated versus α, it is a monotonous increasing function of α for either Fermi-Dirac (F D) or Bose-Einstein (BE) Quantum Statistics (Fig. 6.6). We already know the sign or the value of u(α) for some speciﬁc physical situations : It was noted in § 6.2.2 that the expression of f (ε) for Bose-Einstein statistics implies that α < βε1 , i.e., α < 0 for free bosons, since then ε1 is zero. We

150

Chapter 6. General Properties of the Quantum Statistics

uF D

u± (α)

u+ (α) = uF D u− (α) = uBE uBE constant classical limit αBE

αF D O

N 3 λ Ω th α

Fig. 6.6: Sketch of the function u± (α) and determination of α from the physical data of the considered system.

will see in chapter 9 of this book that, in some situations, α vanishes, thus producing the Bose-Einstein condensation, which is a phase transition. In the case of free fermions at low temperature, it will be seen in the next chapter that µ is larger than ε1 , which is equal to zero, i.e., α > 0. When α tends to minus inﬁnity, the denominator term takes the same expression for both fermions and bosons. The integration is exactly performed : u∓ (α) e

α

+∞

0

√ √ −x α π

1 xe dx = e 2

(6.78)

Considering the expression of the right member of 6.76, this limit does occur if the particle density n = N/Ω tends to zero or the temperature tends to inﬁnity. In the framework of this classical limit, one deduces eα =

N λ3th Ω (2s + 1)

(6.79)

which allows one to express the chemical potential µ, since α = µ/kB T . One

Common Limit of the Quantum Statistics

151

ﬁnds, if the spin degeneracy is neglected, i.e., if one takes 2s + 1 = 1, ñ Å ã ô Ω 2πmkB T 3/2 µ = −kB T ln (6.80) N h2 the very expression obtained in § 4.4.3 for the ideal gas.

6.7.2

Grand Canonical Partition Function of the Ideal Gas

Let us look for the expression of A in the limit where α tends to minus inﬁnity. For both fermions and bosons, one gets the same limit, by developing the Naperian logarithm to ﬁrst order : it is equal to eα−βεk (6.81) A = −kB T k

e−βεk

(6.82)

One recognizes the one-particle partition function z1 = e−βεk

(6.83)

= −kB T e

α

k

k

(its value for the ideal gas was calculated in § 4.4.1), whence A = −kB T eα z1

(6.84)

A = −kB T ln ZG ,

(6.85)

By deﬁnition

thus in the common limit of both Quantum Statistics, i.e., in the classical limit conditions, ZG is equal to ZG = exp(eα z1 ) =

N

eαN

z1N N!

(6.86)

Now the general deﬁnition of the grand canonical partition function ZG , using the N -particle canonical partition function ZcN , is ZG = eαN ZcN (6.87) N

This shows that in the case of identical particles, for sake of consistency, the N -particles canonical partition function must contain the factor 1/N !, the

152

Chapter 6. General Properties of the Quantum Statistics

“memory” of the Quantum Statistics, to take into account the indistinguishability of the particles : ZcN =

1 N z N! 1

(6.88)

This factor ensures the extensivity of the thermodynamical functions, like S, and allows to solve the Gibbs paradox (there is no extra entropy when two volumes of gases of the same density and the same nature are allowed to mix into the same container.) (see § 4.4.3). Notes : The above reasoning allows one to grasp how, in the special case of free particles, one goes from a quantum regime to a classical one. We did not specify to which situation corresponds the room temperature, the most frequent experimental condition. We will see that for metals, for example, it corresponds to a very low temperature regime, of quantum nature (see chapter 7).

Summary of Chapter 6 To account for the occupation conditions of the energy states as stated by the Pauli principle, it is simpler to work in the grand canonical ensemble, in which the number of particles is only given in average : then, at the end of the calculation, the value of the chemical potential is adjusted so that this average number coincides with the real number of particles in the system. As soon as the system is macroscopic, the ﬂuctuations around the average particle number predicted in the grand canonical ensemble are much smaller than the resolution of any measure. The grand partition function gets factorized by energy level and takes a different expression for fermions and for bosons. Remember the distribution, which provides the average number of particles, of chemical potential equal to µ, on the energy level ε at temperature T = 1/kB β : for fermions fF D (ε) =

1 exp β(ε − µ) + 1

(6.89)

fBE (ε) =

1 exp β(ε − µ) − 1

(6.90)

for bosons

For a free particle in a box, we calculated the location of the allowed energy levels. The periodic boundary conditions of Born-Von Kármán (or in an analogous way the stationary-waves conditions) evidence quantized levels, and for a box of macroscopic dimensions (“large volume limit”), the allowed levels are extremely close. One then deﬁnes an energy density of states such that the number of allowed states between ε and ε + dε is equal to dn = D(ε)dε For a free particle in the three-dimensional space D(ε) ∝ 153

(6.91) √

ε.

154

Summary of Chapter 6

In the large volume limit the average values at temperature T of physical parameters are expressed as integrals, which account for the existence of allowed levels through the energy density of states and for their occupation by the particles at this temperature through the distribution term : +∞ D(ε)f (ε) dε (6.92) N= −∞ +∞

U=

εD(ε)f (ε) dε

(6.93)

−∞

We showed how for free particles the Classical Statistics of the ideal gas is the common limit of the Fermi-Dirac and Bose-Einstein statistics when the temperature is very high and/or the density very low. The factor 1/N ! in the canonical partition function of the ideal gas is then the “memory” of the Quantum Statistics properties.

Chapter 7

Free Fermions Properties We now come to the detailed study of the properties of the Fermi-Dirac quantum statistics. First at zero temperature we will study the behavior of free fermions, which shows the consequences of the Pauli principle. Then we will be concerned by electrons, considered as free, and will analyze two of their properties which are temperature-dependent, i.e., their speciﬁc heat and the “ thermionic” emission. In the next chapter we will understand why this free electron model indeed properly describes the properties of the conduction electrons in metals. In fact, in the present chapter we follow the historical order : the properties of electrons in metals, described by this model, were published by Arnold Sommerfeld (electronic speciﬁc heat, Spring 1927) and Wolfang Pauli (electrons paramagnetism, December 1926), shortly after the statement of the Pauli principle at the beginning of 1925.

7.1

7.1.1

Properties of Fermions at Zero Temperature Fermi Distribution, Fermi Energy

The Fermi-Dirac distribution, presented in § 6.5, gives the average value of the occupation number of the one-particle state of energy ε, and is a function of the chemical potential µ. Its expression is ï ò 1 β(ε − µ) 1 f (ε) = β(ε−µ) = 1 − tanh (7.1) 2 2 e +1 155

156

Chapter 7. Free Fermions Properties

1 It is plotted in Fig. 7.1 and takes the value for ε = µ for any value of β and 2 thus of the temperature. This particular point is a center of inversion of the curve, so that the probability of “nonoccupation” for µ− δε, i.e., 1 − f (µ− δε), has the same value as the occupation probability for µ + δε, i.e., f (µ + δε). β At the point ε = µ the tangent slope is − : the lower the temperature, the 4 larger this slope. f 1−f 1

2 kB T

1 − f (ε)

f (ε) 0

µ

ε

Fig. 7.1 : Variation versus energy of f (ε) and 1 − f (ε). In particular when T tends to zero, f (ε) tends to a “step function” (Heaviside function). One usually calls Fermi energy the limit of µ when T tends to zero ; it is noted εF , its existence and its value will be speciﬁed below. Then one obtains at T = 0 : ® f (ε) = 1 if ε < εF (7.2) f (ε) = 0 if ε > εF At T = 0 K the ﬁlled states thus begin at the lowest energies, and end at εF , the last occupied state. The states above εF are empty. This is in agreement with the Pauli principle : indeed in the fundamental state the particles are in their state of minimum energy but two fermions from the same physical system cannot occupy the same quantum state, whence the requirement to climb up in energy until the particles all located. In the framework of the large volume limit, to which this course is limited, the total number N of particles in the system satisﬁes ∞ D(ε)f (ε)dε (7.3) N= 0

where D(ε) in the one-particle density of states in energy. This condition determines the µ value.

Properties of Fermions at Zero Temperature

157

At zero temperature, the chemical potential, then called the “Fermi energy εF ,”, is thus deduced from the condition that εF N= D(ε)dε (7.4) 0

The number N corresponds to the hatched area of Fig. 7.2.

D(ε)

f (ε)

T = OK

0

εF

ε

Fig. 7.2: Density of states D(ε) of three-dimensional free particles and determination of the Fermi energy at zero temperature. The hatched area under the curve provides the number N of particles in the system. For a three-dimensional system of free particles, with a spin s = 1/2, from (6.64) √ Ω (2m)3/2 ε 2π 2 3 D(ε) = 0

D(ε) =

for ε > 0 for

(7.5)

ε≤0

The integral (7.4) is then equal to N=

2 Ω 3/2 (2m)3/2 εF 3 2π 2 3

(7.6)

Å ã N 2/3 2 3π 2 2m Ω

(7.7)

that is, εF =

This expression leads to several comments :

158

Chapter 7. Free Fermions Properties

– N and Ω are extensive, thus εF is intensive ; – εF only depends on the density N/Ω ; – the factor of 2 /2m in (7.7) is equal to kF2 , where kF is the Fermi wave vector : in the k-space the surface corresponding to the constant energy εF , called “Fermi surface,” contains all the ﬁlled states. In the case of free particles, it is a sphere, the “Fermi sphere,” of radius kF . One can √ also deﬁne the momentum pF = kF = 2mεF . Then expression (7.6) can be obviously expressed versus kF or pF for spin-1/2 particles : N=

2Ω 4 3 2Ω 4π 3 πk = 3 p (2π)3 3 F h 3 F

(7.8)

p)), the one-particle density The ﬁrst factor is equal to Dk (k) (resp. Dp ( of states in the wave vector- (resp. momentum-) space. Expression (7.8) is also directly obtained by writing : N= 0

kF

Dk (k)d3k .

– εF corresponds, for metals, to energies of the order of a few eV. Indeed, let us assume now (see next chapter) that the only electrons to be considered are those of the last incomplete shell, which participate in the conduction. Take the example of copper, of atomic mass 63.5 g and density 9 g/cm3 . There are thus 9N /63.5 = 0.14N copper atoms per cm3 , N being the Avogadro number. Since one electron per copper atom participates in conductivity, the number of such electrons per cm3 is of 0.14N = 8.4 × 1022 /cm3 i.e., 8.4 × 1028 / m3 . This corresponds to a wave vector kF = 1.4 × 1010 m−1 , and to a velocity at the Fermi level kF of 1.5 × 106 m/sec. One deduces εF = 7 eV. vF = m Remember that the thermal energy at 300 K, kB T , is close to 25 meV, so that the Fermi energy εF corresponds to a temperature TF such that kB TF = εF , in the 104 or 105 K range according to the metal (for copper 80, 000 K) !

7.1.2

Internal Energy and Pressure of a Fermi Gas at Zero Temperature

The internal energy at zero temperature is given by εF εD(ε)dε U= 0

(7.9)

Properties of Fermions at Zero Temperature

159

For the three-dimensional density of states (7.5), U is equal to : U=

2 Ω 5/2 (2m)3/2 εF 5 2π 2 3

(7.10)

i.e., by expressing N using (7.6), U= Besides,

3 N εF 5 Å

∂U P =− ∂Ω

(7.11)

ã (7.12) N

From (7.10) and (7.7), U varies like N 5/3 Ω−2/3 , so that at constant particles number N P =

2U 3Ω

(7.13)

For copper, one obtains P = 3.8 × 1010 N/m2 = 3.7 × 105 atmospheres. This very large pressure, exerted by the particles on the walls of the solid which contains them, exists even at zero temperature. Although the state equation (7.13) is identical in the cases of fermions and of the ideal gas, the two situations are very diﬀerent : for a classical gas, when the temperature tends to zero, both the internal energy and the pressure vanish. For fermions, the Pauli principle dictates that levels of nonzero energies are ﬁlled. The fact that electrons in the same orbital state cannot be more than a pair, and of opposite spins, leads to their mutual repulsion and to a large average momentum, at the origin of this high pressure.

7.1.3

Magnetic Properties. Pauli Paramagnetism

A gas of electrons, each of them carrying a spin magnetic moment equal to If N+ the Bohr magneton µ B , is submitted to an external magnetic ﬁeld B. equal to +µB , is the number of electrons with a moment projection along B N− that of moment −µB , the gas total magnetic moment (or magnetization) is along B M = µB (N+ − N− )

(7.14)

The density of states Dk (k) is the same as for free electrons. Since the potential energy of an electron depends on the direction of its magnetic moment, the total energy of an electron is ε± =

2k 2 ∓ µB B 2m

(7.15)

160

Chapter 7. Free Fermions Properties

The densities of states D+ (ε) and D− (ε) now diﬀer for the two magnetic moment orientations, whereas in the absence of magnetic ﬁeld, D+ (ε) = D− (ε) = D(ε)/2 : here ® D+ (ε) = 12 D(ε + µB B) (7.16) D− (ε) = 12 D(ε − µB B) Besides, the populations of electrons with the two orientations of magnetic moment are in equilibrium, thus they have the same chemical potential. The solution of this problem is a standard exercise of Statistical Physics. Here we just give the result, in order to comment upon it : one obtains the magnetization M = µB · µB B · D(εF )

(7.17)

This expresses that the moments with the opposite orientations compensate, except for a small energy range of width µB B around the Fermi energy. The 1 M corresponding (Pauli) susceptibility, deﬁned by lim , is equal to µ0 Ω B B→0 χ=

1 3 N µ2B kB T 1 3 N µ2B · = µ0 2 Ω εF µ0 2 Ω kB T εF

(7.18)

with µ0 = 4π × 10−7 . It is reduced with respect to the classical result (Curie law) by a factor of the order of kB T /εF , which expresses that only a very small fraction of the electrons participate in the eﬀect (in other words, only a few electrons can reverse their magnetic moment to have it directed along the external ﬁeld).

7.2

7.2.1

Properties of Fermions at Non-Zero Temperature Temperature Ranges and Chemical Potential Variation

At nonvanishing temperature, one again obtains the chemical potential µ position from expression (7.3), which links µ to the total number of particles : ∞ D(ε)f (ε)dε N= 0 ∞ 1 = D(ε) β(ε−µ) dε (7.3) e +1 0

Properties of Fermions at Non-Zero Temperature

161

D(ε) f (ε) T TF

f (ε)D(ε)

0

µ

εF

ε

Fig. 7.3 : Determination of the chemical potential µ for T TF .

Relation (7.3) implies that µ varies with T . Fig. 7.3 qualitatively shows for kB T εF the properties which will be calculated in Appendix 7.1. Indeed the area under the curve representing the product D(ε)f (ε) gives the number of particles. At T = 0 K, one has to consider the area under D(ε) up to the energy εF ; at small T , a few states above µ are ﬁlled and an equivalent number below µ are empty. Since the function D(ε) is increasing, to maintain an area equal to N, µ has to be smaller than εF , of an amount proportional to D (µ). The exact relation allowing one to deduce µ(T ) for the two Quantum Statistics was given in (6.74) : recalling that α = βµ, and substituting (7.5) into (7.3) one obtains for fermions √ ∞ Å ã3/2 xdx 2 2 N = 2π · (7.19) exp(x − α) + 1 Ω 2mkB T 0 Å ã 2 εF 3/2 = (7.20) 3 kB T In the integral we deﬁned βε = x. Thus α is related to the quantity εF /kB T = TF /T . – When T vanishes, the integral (7.19) tends to inﬁnity : the exponential in the denominator no longer plays any role since α tends to inﬁnity (εF is ﬁnite and β tends to inﬁnity). The right member also tends to

162

Chapter 7. Free Fermions Properties

inﬁnity. One thus has to return to equation (7.4). – We have just studied the low temperature regime T TF , which takes place in metals at room temperature (TF ∼ 104 K) and corresponds to Fig. 7.3. In astrophysics, the “white dwarfs” enter in this framework. They are stars at the end of their existence, which have already burnt all their nuclear fuel and therefore have shrunk. Their size is of the order of a planet (radius of about 5000 km) for a mass of 2 × 1030 kg (of the order of that of the sun), their temperature T of 107 K. Their high density in electrons determines εF . TF being in the 109 K range, the low temperature regime is achieved ! This regime for which the Quantum Statistics apply is called the “degenerate” (in opposition to “classical”) regime : as already mentioned about Pauli paramagnetism, and will be seen again below, the “active” particles all have an energy close to µ, this is almost as if this level was degenerate in the Quantum Mechanics meaning. – On the other hand, when T TF , the integral (7.19) vanishes. This is realized either at low density (εF small), or at given density when T tends to inﬁnity. Then e−α tends to inﬁnity, and since α = βµ with β tending to zero, this implies that µ tends to −∞. One then ﬁnds the Maxwell-Boltzmann Classical Statistics, as discussed in § 6.7.2. – Between these two extreme situations, µ thus decreases from εF to −∞ when the ratio T /TF increases. When studying systems at room temperature, the regime is degenerate in most cases. For properties mainly depending on the Pauli principle, one can consider then to be at zero temperature. This is the situation when ﬁlling the electronic shells in atoms according to the “Aufbau (construction) principle”, widely used in Chemistry courses : the Quantum Mechanics problem is ﬁrst solved in order to ﬁnd the one-electron energy levels of the considered atom. The solutions ranked in increasing energies are the states 1s, 2s, 2p, 3s, 3p, 4s, 3d, etc. (except some particular cases for some transition metals). These levels are then ﬁlled with electrons, shell after shell, taking into account the level degeneracy (K shell, n = 1, containing at most two electrons, L shell of n = 2, with at most eight electrons, etc.). Since the energy splittings between levels are of the order of a keV for the deep shells and of an eV for the valence shell, the thermal energy remains negligible with respect to the energies coming into play and it is as if the temperature were zero. On the other hand, if one is interested in temperature-dependent properties, like the speciﬁc heat or the entropy, which are derivatives versus temperature of thermodynamical functions, more exact expressions are required and at low temperature one will proceed through limited developments versus the variable T /TF .

Properties of Fermions at Non-Zero Temperature

7.2.2

163

Speciﬁc Heat of Fermions

The very small contribution of electrons to a metal speciﬁc heat, which follow the Dulong and Petit law (see Ch. 2, § (2.4.5)) only introducing the vibrations of the lattice atoms, presented a theoretical challenge in the 1910s. This lead to a questioning of the electron conduction model, as proposed by Paul Drude in 1900. Å ã ∂U of a fermions assembly, one To calculate the speciﬁc heat Cv = ∂T N,Ω needs to express the temperature dependence of the internal energy. One can already make an estimate of Cv from the observation of the shape of the Fermi distribution : the only electrons which can absorb energy are those occupying the states in a region of a few kB T width around the chemical potential, and each of them will increase its energy by an amount of the order kB T to reach an empty state. The number of concerned electrons is thus ∼ kB T · D(µ) and for a threekB T dimensional system at low temperature (

1, so that µ and εF are εF close), µ N∼ D(ε)dε ∝ µ3/2 , (7.21) 0

3N dN = = D(µ) i.e., dµ 2µ

(7.22)

The total absorbed energy is thus of the order of Å ã 3N U ∼ kB T · · kB T 2 µ Consequently, Cv =

dU ∼ N kB dT

Å

kB T µ

(7.23)

ã (7.24)

These considerations evidence a reduction of the value of Cv with respect to the classical value N kB in a ratio of the order kB T /µ 1, because of the very small number of concerned electrons. The exact approach refers to a general method of calculation of the integrals in which the Fermi distribution enters : the Sommerfeld development. The complete calculation is given in Appendix 7.1. Here we just indicate its principle. At a temperature T small with respect to TF , one looks for a limited development in T /TF of the fermions internal energy U , which depends of T , and of

164

Chapter 7. Free Fermions Properties

the chemical Å ã potential, itself temperature-dependent. The searched quantity dU is then equal to Cv = dT N,Ω Cv =

3 π2 N kB · 2 3

Å

kB T εF

ã (7.25)

which is indeed of the form (7.24), as far as µ is not very diﬀerent from εF (see Appendix 7.1). The ﬁrst factor in (7.25) is the speciﬁc heat of a monoatomic ideal gas ; the last factor is much smaller than unity in the development conditions, it ranges between 10−3 and 10−2 at 300 K (for copper it is 3.6 × 10−3 ) and expresses the small number of eﬀective electrons. The electron’s speciﬁc heat in metals is measured in experiments performed at temperatures in the kelvin range. In fact the measured total speciﬁc heat includes the contributions of both the electrons and the lattice Cvtotal = Cvel + Cvlattice

(7.26)

The contribution Cvlattice of the lattice vibrations to the solid speciﬁc heat was studied in Ch. 2, § 2.4.5. The Debye model was mentioned, which correctly interprets the low temperature results by predicting that Cvlattice is proportional to T 3 , as observed in non-metallic materials. The lower the temperature, the less this contribution will overcome that of the electrons. The predicted variation versus T for Cvtotal is thus given by Cvtotal = γT + aT 3

(7.27)

Cvtotal

= γ + aT 2

(7.28)

π 2 kB 2 εF

(7.29)

T with, from (7.25),

γ = N kB

From the measurement of γ one deduces a value of εF of 5.3 eV for copper whereas in § 7.1.1 we calculated 7 eV from the value of the electrons concentration, using (7.7). In the case of silver the agreement is much better since the experiment gives εF = 5.8 eV, close to the prediction of (7.7). The possible discrepancies arise from the fact that the mass m used in the theoretical expression of the Fermi energy was that of the free electron. Now a conduction electron is submitted to interactions with the lattice ions, with the solid vibrations and the other electrons. Consequently, its motion is altered,

Properties of Fermions at Non-Zero Temperature

165

which can be expressed by attributing it an eﬀective mass m∗ . This concept will be explained in § 8.2.3 and § 8.3.3. Deﬁning εF =

ã2/3 Å 2 2N , 3π 2m∗ V

(7.30)

one deduces from these experiments a mass m∗ = 1.32m for copper, m∗ = 1.1m for silver, where m is the free electron mass. Note that the fermion entropy can be easily deduced from the development of the Appendix : we saw that A = −P Ω (see §Å3.5.3), ã so that, according to 2 ∂A (7.13), A = − U . Besides, from (3.5.3), S = − . One deduces 3 ∂T µ,Ω S = N kB

π 2 kB T = Cv 2 εF

(7.31)

The electrons entropy thus satisﬁes the third law of Thermodynamics since it vanishes with T .

7.2.3

Thermionic Emission ε

0

V εF

f (ε) 1

0

Fig. 7.4: Potential well for the electrons of a metal. In order to extract an electron at zero temperature, one illuminates the solid with a photon of energy higher than V − εF . The walls of the “potential box” which conﬁnes the electrons in a solid are not really inﬁnitely high. The photoelectrical eﬀect, evidenced by Heinrich

166

Chapter 7. Free Fermions Properties

Hertz (1887), and interpreted by Albert Einstein (this was the argument for Einstein’s Nobel prize in 1921), allows one to extract electrons from a solid, at room temperature, by illumination with a light of high enough energy : for zinc, a radiation in the near ultraviolet is required (Fig. 7.4). The energy’s origin is here the state of a free electron with zero kinetic energy, i.e., with such a convention the Fermi level has a negative energy. The potential barrier is equal to V . To extract an electron at zero temperature, a photon must bring it at least V − εF , an energy usually in the ultraviolet range. When heating this metal to a high temperature T , the Fermi distribution is spreading toward little negative or even positive energies, so that a few electrons are then able to pass the potential barrier and become free (Fig. 7.5). One utilizes tungsten heated to about 3000 K to make the ﬁlaments of the electrons cannons of TV tubes or electronic microscopes, as its fusion temperature is particularly high (3700 K). ε

0

V µ

f (ε)

1/2

Fig. 7.5: Filling of the electronic states in a metal heated to a high temperature T . One way to calculate the thermal current consists in expressing that a particle escapes the metal only if it can overcome the potential barrier, i.e. if its kinetic energy in the direction perpendicular to the surface, taken as z axis, is large enough. It should satisfy p2 p2 ≥ z ≥V (7.32) 2m 2m We know that the electronic energies, and thus V and V − µ, are of the order of a few eV : whatever the ﬁlament temperature, the system is in a regime

Properties of Fermions at Non-Zero Temperature

167

where kB T remains small with respect to the concerned energies, and where kB T

1. Consequently, µ is still close to εF , and in the Fermi distribution εF f (ε) = exp β

Ä

1 p2 2m

ä −µ +1

(7.33)

the denominator exponent is large for the electrons likely to be emitted. The Fermi distribution can then be approximated by a Maxwell-Boltzmann distribution p2

f (ε) eβµ e−β 2m

(7.34)

There remains to estimate the number ∆n of electrons moving toward the surface ∆S during a time ∆t. This is a standard problem of kinetic theory of gases (the detailed method of calculation is analogous to the kinetic calculation of the pressure, § 4.2.2). One obtains a total current jz = ∆n/∆S∆t equal to jz =

2 ∆n = 3 2πm(kB T )2 e−β(V −µ) ∆S∆t h

(7.35)

This expression shows the very fast variation of the thermal emission current with the temperature T , in T 2 e−(V −µ)/kB T . It predicts that a 0.3 × 0.3 mm2 tungsten tip emits a current of 15 mA at 3000 K and of only 0.8µA at 2000 K. Practically, to obtain a permanent emission current, one must apply an external electric ﬁeld : in the absence of such a ﬁeld, the ﬁrst electrons emitted into vacuum create a space charge, associated with an antagonistic electric ﬁeld, which prevents any further emission. Another way of calculating the emitted current consists in considering that in the TV tube, in the absence of applied electrical potential, an equilibrium sets up at temperature T between the tungsten electrons which try to escape the metal and those outside the metal that try to enter it. The equilibrium between electrons in both phases is expressed through the equality of their chemical potentials. Now, outside the solid the electron concentration is very small, so that one deals with a classical gas which follows the Maxwell-Boltzmann statistics. The electron current toward the solid is calculated like for an ideal gas. When an electric ﬁeld is applied, this current vanishes but the emission current, which is its opposite, remains. Note that, in order to be rigorous, all these calculations should take into account the energy levels of the metal and its surface, which would be very

168

Chapter 7. Free Fermions Properties

complicated ! Anyhow, the above estimates qualitatively describe the eﬀect and lead to correct orders of magnitude. On this example of the thermionic emission we have shown a practical application, in which electrons follow the high temperature limit of the Fermi-Dirac statistics.

Summary of Chapter 7 The properties of free fermions are governed by the Fermi-Dirac distribution : fF D (ε) =

1 exp β(ε − µ) + 1

At zero temperature the fermions occupy all the levels of energy lower or equal to εF , the Fermi energy, which is deﬁned as the chemical potential µ for T = 0 K. The number N of particles of the system and the Fermi energy are linked at T = 0 K through εF N= D(ε)dε 0

This situation, related to the Pauli principle, induces large values of the internal energy and of the pressure of a system of N fermions. An excitation, of thermal or magnetic origin, of a fermions system, only aﬀects the states of energy close to µ or εF . Consequently, the fermions’ magnetic susceptibility and speciﬁc heat are reduced with respect to their values in a classical system. In the case of the speciﬁc heat, the reduction is a ratio of the order of kB T /µ, where µ is the chemical potential, of energy close to εF . The magnetic susceptibility and the speciﬁc heat of electrons in metals are well described in the framework of the free electrons model developed in this chapter.

169

Appendix 7.1 Calculation of Cv Through the Sommerfeld Development ã dU of fermions is deduced from the expression dT N,Ω of the internal energy U (T ) versus temperature. Å

The speciﬁc heat Cv =

We have therefore to calculate, at nonvanishing temperature, ∞ f (ε)D(ε)dε N= 0 ∞ U= εf (ε)D(ε)dε

(7.36) (7.37)

0

that is, more generally, g(ε) =

+∞

+∞

g(ε)D(ε)f (ε)dε =

ϕ(ε)f (ε)dε

−∞

(7.38)

−∞

where ϕ(ε) is assumed to vary like a power law of ε for ε > εminimum, to be null for smaller values of ε. We know that, owing to the shape of the Fermi distribution, εF lim g(ε) = ϕ(ε)dε T →0

We write the diﬀerence

+∞

∆= −∞

(7.39)

−∞

ϕ(ε)f (ε)dε −

µ

ϕ(ε)dε

(7.40)

−∞

which is small when kB T µ (βµ 1). This diﬀerence introduces ϕ (µ) : if ϕ (µ) is zero, because of the symmetry of f (ε) around µ, the contributions on 171

172

Appendix 7.1

both sides of µ compensate. Besides, we understood in § 7.2.1 that, to maintain the same N when the temperature increases, µ gets closer to the origin, for a three-dimensional system for which ϕ (µ) > 0 when D(ε) is diﬀerent from √ zero (ϕ(ε) ∝ ε). Let us write ∆ explicitly : µ ∆= dεϕ(ε)f (ε) − −∞ µ

= −∞

1 eβ(ε−µ)

∞

dεϕ(ε) +

−∞

Å dεϕ(ε)

µ

+1

dεϕ(ε)f (ε)

(7.41)

µ

ã −1 +

∞

dεϕ(ε)

µ

1 eβ(ε−µ)

+1

(7.42)

Here a symmetrical part is played by either the empty states below µ [the probability for a state to be empty is 1 − f (ε)], or the ﬁlled states above µ, all in small numbers if the temperature is low. We take β(ε − µ) = x, βdε = dx ∞ 0 Å Å ã ã x x dx dx 1 1 ∆=− ϕ µ+ ϕ µ+ + −x β β 1 + e β β 1 + ex −∞ 0 ∞ ∞ Å Å ã ã x 1 x 1 dx dx ϕ µ− ϕ µ+ + =− x β β 1 + e β β 1 + ex ∞0 ï Å ã Å 0 ãò x x dx 1 ϕ µ+ − ϕ µ− = x β 1 + e β β 0 ñ ô Å ã 1 ∞ dx x ϕ(3) (µ) x 3 = · 2 ϕ (µ) + + ... β 0 1 + ex β 3! β =

∞ ϕ(2n+1) (µ) In β 2n+2 n=0

(7.43) (7.44) (7.45) (7.46) (7.47)

with In =

2 (2n + 1)!

∞

dx 0

x2n+1 ex + 1

(7.48)

The integrals In are deduced from the Riemann ζ functions (see the section “A few useful formulae”). In particular at low temperature the principal term π2 of ∆ is equal to (kB T )2 ϕ (µ). 6 Let us return to the calculation of the speciﬁc heat Cv . For the internal energy of nonrelativistic particles, ϕ(ε) = ε · Kε1/2 = Kε3/2 for ε > 0, ϕ(ε) = 0 for ε < 0. Then µ π2 U (T ) = (kB T )2 (Kε3/2 )µ + O(T 4 ) Kε3/2 dε + (7.49) 6 0

Appendix 7.1

173

but µ depends on T since µ εF π2 1/2 2 1/2 4 N= (kB T ) (Kε )µ + O(T ) = Kε dε + Kε1/2 dε 6 0 0

(7.50)

expresses the conservation of the particles number. One deduces from (7.50) : ñ µ(T ) = εF

π2 1− 12

Å

kB T εF

ã2 ô + O(T 4 )

(7.51)

Replacing the expression of µ(T ) into U (T ), one gets : 5/2 π

U (T ) − U (T = 0) KεF

2

6

Å

kB T εF

ã2 = N εF

π2 4

Å

kB T εF

ã2 (7.52)

One deduces Cv : 3 dU π2 = N kB · Cv = dT 2 3

Å

kB T εF

ã (7.53)

Chapter 8

Elements of Bands Theory and Crystal Conductivity The previous chapter on free fermions allowed us to interpret some properties of metals. We considered that each electron was conﬁned within the solid by potential barriers, but did not introduce any speciﬁc description of its interactions with the lattice ions or with the other electrons. This model is not very realistic, in particular it cannot explain why some solids are conducting the electrical current and others are not : at room temperature, the conductivity of copper is almost 108 siemens (Ω−1 · m−1 ) whereas that of quartz is of 10−15 S, and of teﬂon 10−16 S. This physical parameter varies on 24 orders of magnitude according to the material, which is absolutely exceptional in nature. Thus we are going to reconsider the study of electrons in solids from the start and to ﬁrst address the characteristics of a solid and of its physical description. As we now know, the ﬁrst problem to solve is a quantum mechanical one, and we will rely on your previous courses. In particular we will be concerned by a periodic solid, that is, a crystal (§ 8.1). We will see that the electron eigenstates are regrouped into energy bands, separated by energy gaps (§ 8.2 and § 8.3). We will have to introduce here very simplifying hypotheses : in Solid State Physics courses, more general assumptions are used to describe electronic states of crystals. Using Statistical Physics (§ 8.4) we will see how electrons are ﬁlling energy bands. We will understand why some materials are conductors, other ones insulators. We will go more into details on the statistics of semiconductors,

175

176

Chapter 8. Elements of Bands Theory and Crystal Conductivity

of high importance in our everyday life, and will end by a brief description of the principles of a few devices based on semiconductors (§ 8.5).

8.1

What is a Solid, a Crystal ?

A solid is a dense system with a speciﬁc shape of its own. This deﬁnition expresses mechanical properties ; but one can also refer to the solid’s visual aspect, shiny in the case of metals, transparent in the case of glass, colored for painting pigments. These optical properties, as well as the electrical conduction and the magnetic properties, are due to the electrons. In a macroscopic solid with a volume in the cm3 range, the numbers of electrons and nuclei present are of the order of the Avogadro number N = 6.02 × 1023 . The hamiltonian describing this solid contains for each particle, either electron or nucleus, its kinetic energy, and its potential energy arising from its Coulombic interaction with all the other charged particles. It is easily guessed that the exact solution will be impossible to get without approximations (and still very diﬃcult to obtain using them !). One ﬁrst notices that nuclei are much heavier than electrons and, consequently, much less mobile : either the nuclei motion will be neglected by considering that their positions remain ﬁxed, or the small oscillations of these nuclei around their equilibrium positions (solid vibrations) will be treated seˆ will parately from the electron motions, so that the electron hamiltonian H ˆ will then be given by not include the nuclei kinetic energy. H ˆ = H

NZ N NZ N Z ˆ2 pi 1 1 Ze2 1 e2 − + n | 2 4πε0 2m 4πε0 i=1 n=1 |ri − R |r − rj | i=1 i,j=1 i

(8.1)

n and the where Z is the nuclear charge ; the nth nucleus is located at R electron i at ri . The second term expresses the Coulombian interaction of all the electrons with all the nuclei, the third one the repulsion among electrons. Besides, we will assume that the solid is a perfect crystal, that is, its atoms are n of a three-dimensional periodic lattice. This regularly located on the sites R will simplify the solution of the Quantum Mechanical problem, but many of our conclusions will also apply to disordered solids. ˆ depends of their probabilities of The repulsion term between electrons in H presence, in turn given by the solution of the hamiltonian : this is a diﬃcult self-consistent problem. The Hartree method, which averages the electronic repulsion, and the Hartree-Fock method, which takes into account the wave function antisymmetrization as required by the Pauli principle, are used to

The Eigenstates for the Chosen Model

177

solve this problem by iteration, but they lead to complicated calculations. Phenomena like ferromagnetism, superconductivity, Mott insulators, mixedvalence systems, are based on the interaction between electrons ; however in most cases one can account of the repulsion between electrons in a crystal through an average one-electron periodic term and we will limit ourselves to this situation in the present course. Then it will be possible to split the ˆ into one-electron terms, this is the independent-electron aphamiltonian H proximation : ˆ = H

N

ˆi h

(8.2)

i=1

with ˆ2 p ˆ hi = i + V(ri ) 2m

(8.3)

The potential V(ri ) has the lattice periodicity, it takes into account the attractions and repulsions on electron i and is globally attractive. The eigenvalues ˆ are the sums of the eigenenergies of the various electrons. of H ˆ i . We will Our task will consist in solving the one-electron hamiltonian h introduce an additional simpliﬁcation by assuming that the lattice is onedimensional, i.e., V(ri ) = V(xi ). This will allow us to extract the essential characteristics, at the cost of simpler calculations. In the “electron in the box” model that we have used up to now, V(xi ) only expressed the potential barriers limiting the solid, and thus leveled out the potential of each atom. Here we will introduce the attractive potential on each lattice site. The energy levels will derive from atomic levels, rather than from the free electron states which were the solutions in the “box” case.

8.2 8.2.1

The Eigenstates for the Chosen Model Recall : the Double Potential Well

First consider an electron on an isolated atom, located at the origin. Its hamiltonian ˆ h is given by 2 ˆ = pˆ + V (x) h 2m

where V is the even Coulombian potential of the nucleus (Fig. 8.1).

(8.4)

178

Chapter 8. Elements of Bands Theory and Crystal Conductivity

One of its eigenstate is such that : ˆ hϕ(x) = ε0 ϕ(x)

(8.5)

ϕ(x) is an atomic orbital, of s-type for example, which decreases from the origin with a characteristic distance λ. We will assume that the other orbitals do not have to be considered in this problem, since they are very distant in energy from ε0 . V

λ

|ϕ|2

x

Fig. 8.1: Coulombian potential of an isolated ion and probability of presence of an electron on this ion. Let us add another nucleus, that we locate at x = d (Fig. 8.2). The hamiltonian, still for a single electron now sharing its presence between the two ions, 2 ˆ = pˆ + V (x) + V (x − d) h 2m

(8.6)

In the case of two 1s-orbital initial states, this corresponds to the description of the H+ 2 ion (two protons, one electron). One follows the standard method on the double potential well of the Quantum Mechanics courses and deﬁnes |ϕL and |ϕR , the atomic eigenstates respectively located on the left nucleus (at x = 0 ) and the right nucleus (at x = d) : ⎧ñ ô ⎪ pˆ2 ⎪ ⎪ + V (x) |ϕL = ε0 |ϕL ⎪ ⎪ ⎨ 2m (8.7) ⎪ ï 2 ò ⎪ ⎪ pˆ ⎪ ⎪ + V (x − d) |ϕR = ε0 |ϕR ⎩ 2m

The Eigenstates for the Chosen Model

179

V

d

0

x ε0 + A ε0 − A

Fig. 8.2: Potential due to two ions of the same nature, distant of d and corresponding bound states. Owing to the symmetry of the potential in (8.6) with respect to the point ˆ can be chosen either symmetrical x = d/2, whatever d the eigenfunctions of h or antisymmetrical with respect to this mid-point. However, if d is very large ˆ has two with respect to λ, the coupling between the two sites is very weak, h eigenvalues practically equal to ε0 and if at the initial time the electron is on one of the nuclei, for example the left one L, it will need a time overcoming the observation possibilities to move to the right one R. One can then consider that the atomic states (8.7) are the eigenstates of the problem (although strictly they are not). On the other hand, you learnt in Quantum Mechanics that if d is not too large with respect to λ, the L and R wells are coupled by the tunnel eﬀect, which is expressed through the coupling matrix element ˆ L , −A = ϕR |H|ϕ

A>0

(8.8)

which decreases versus d approximately like exp(−d/λ). ˆ are equal The energy degeneracy is then lifted : the two new eigenvalues of h 1 to ε0 − A, corresponding to the symmetrical wave function |ψS = √ (|ϕL + 2 1 |ϕR ), and ε0 + A, of antisymmetrical wave function |ψA = √ (|ϕL − |ϕR ) 2 (Fig. 8.2). These results are valid for a weak enough coupling (A |ε0 |). For a stronger coupling (d ∼ λ) one should account for the overlap of the two atomic functions, expressed by ϕL |ϕR diﬀerent from zero. Anyway the eigenstates are delocalized between both wells. For H+ 2 they correspond to the bonding and antibonding states of the chemical bound.

180

8.2.2

Chapter 8. Elements of Bands Theory and Crystal Conductivity

Electron on an Inﬁnite and Periodic Chain

In the independent electrons model, the hamiltonian describing a single electron in the crystal is given by +∞ 2 ˆ crystal = pˆ + V (x − xn ) h 2m n=−∞

(8.9)

where xn = nd, is the location of the nth nucleus. The potentials sum is periodic and plays the role of V(ri ) in formula (8.3) (Fig. 8.3). V

0

d

2d

nd x

Fig. 8.3 : Potential periodic of a crystal. One is looking for the stationary solution, such that ˆ crystal ψ(x) = εψ(x) h

(8.10)

If the nuclei are far enough apart, there is no coupling : although the eigenstates are delocalized, if at the initial time the electron is on a given nucleus, an inﬁnite time would be required for this electron to move to the neighboring nucleus. The state of energy ε = ε0 is N times degenerate. On the other hand, in the presence of an even weak coupling, the eigenstates are delocalized. We are going to assume that the coupling is weak and will note : ˆn + ˆ crystal = h V (x − n d) (8.11) h n =n

to show the atomic hamiltonian on the speciﬁc site n. To solve the eigenvalue problem, by analogy with the case of the double potential well, we will work in a basis of localized states, which are not the eigenstates of the considered inﬁnite periodic chain. Now every site plays a similar role, the problem is invariant under the translations which change x into x + nd = x + xn . The general basis state is written |n, it corresponds to the wave function φn (x) located around xn and is deduced through the translation xn from the function centered at the origin, i.e., φn (x) = φ0 (x − xn ). These states are built through orthogonalization from the atomic functions ϕ(x − xn ) in order to constitute an orthonormal basis : n |n = δnn (this is

The Eigenstates for the Chosen Model

181

the so-called “Hückel model”). The functions φn (x) depend on the coupling. When the atoms are very distant, φn (x) is very close to ϕ(x − xn ), the atomic wave function. In a weak coupling situation, where the tunnel eﬀect only takes place between ﬁrst neighbors, φn (x) is still not much diﬀerent from the atomic function. One looks for a stationary solution of the form |ψ = cn |n

(8.12)

n

One needs to ﬁnd the coeﬃcients cn , expected to all have the same norm, since every site plays the same role : ˆ hcrist

+∞

+∞ cn |n = ε cn |n

n=−∞

(8.13)

n=−∞

One proceeds by analogy with the solution of the double-well potential problem : one assumes that the hamiltonian only couples the ﬁrst neighboring sites. Then the coupling term is given by ˆ crist |n + 1 = n − 1|h ˆ crist |n = −A n|h

(8.14)

For atomic s-type functions ϕ(x) A is negative. By multiplying (8.13) by the bra n|, one obtains ˆ cryst |n − 1 + cn n|h ˆ cryst |n + cn+1 n|h ˆ cryst |n + 1 = cn ε cn−1 n|h Now ˆ cryst |n = n|h ˆ n |n + n| n|h

(8.15)

V (x − n d)|n

n =n

(8.16)

= ε0 + α ε0 The term α, which is the integral of the product of |φn (x)|2 by the sum of potentials of sites distinct from xn , is very small in weak coupling conditions [it varies in exp(−2d/λ), whereas A is in exp(−d/λ)]. It will be neglected with respect to ε0 . The coeﬃcients cn , cn−1 and cn+1 are then related through −cn−1 A + cn ε0 − cn+1 A = cn ε

(8.17)

One obtains analogous equations by multiplying (8.13) by the other bras, whence ﬁnally the set of coupled equations : ⎧ ⎪ ⎨. . . . . . (8.18) −cn−1 A + cn ε0 − cn+1 A = cn ε ⎪ ⎩ − cn A + cn+1 ε0 − cn+2 A = cn+1 ε

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8.2.3

Energy Bands and Bloch Functions

The system (8.18) assumes a nonzero solution if one chooses cn as a phase : cn = exp(ik · nd)

(8.19)

where k, homogeneous to a reciprocal length, is a wave vector and nd the coordinate of the considered site. By substituting (8.19) into (8.17) one obtains the dispersion relation, or dispersion law, linking ε and k : ε(k) = ε0 − 2A cos kd

(8.20)

ε(k) is an even function, periodic in k, which describes the totality of its values, i.e. the interval [ε0 − 2A, ε0 + 2A], when k varies from −π/d to π/d (Fig. 8.4). The domain of accessible energies ε constitutes an allowed band ; a value of ε chosen outside of the interval [ε0 − 2A, ε0 + 2A] does not correspond to a kvalue, it belongs to a forbidden domain. The range −π/d ≤ k < π/d is called “the ﬁrst Brillouin zone,” a concept introduced in 1930 by Léon Brillouin, a French physicist ; it is suﬃcient to study the physics of the problem. Note that ε(k)

−π/d

π/d

k

ε0 + 2 A ε0

4A

ε0 − 2 A

Fig. 8.4 : Dispersion relation in the crystal. in the neighborhood of k = 0, ε(k) ε0 − 2A + Ad2 k 2 + . . . which looks like the free electron dispersion relation, if an eﬀective mass m∗ is deﬁned such that 2 /2m∗ = Ad2 . This mass also expresses the smaller or larger diﬃculty for an electron to move under the application of an external electric ﬁeld, due to the interactions inside the solid it is subjected to. The wave function associated with ε(k) is given by ψk (x) =

+∞ n=−∞

exp iknd · φ0 (x − nd)

(8.21)

The Eigenstates for the Chosen Model

183

This expression can be transformed into ψk (x) = exp ikx ·

+∞

exp −ik(x − nd) · φ0 (x − nd)

(8.22)

n=−∞

which is the product of a plane wave by a periodic function uk (x), of period d : indeed, for n integer, uk (x + n d) =

+∞

exp −ik(x + n d − nd) · φ0 (x + n d − nd)

n=−∞

=

+∞

(8.23)

exp −ik(x − n d) · φ0 (x − n d) = uk (x)

n =−∞

This type of solution of the Schroedinger stationary equation ψk (x) = eikx · uk (x)

(8.24)

can be generalized, for a three-dimensional periodic potential V (r), into

ψk (r) = eik·r · uk

(8.25)

This is the so-called Bloch function, from the name of Felix Bloch, who, in his doctoral thesis prepared in Leipzig under the supervision of Werner Heisenberg and defended in July 1928, was the ﬁrst to obtain the expression of a crystal wave function, using a method analogous to the one followed here. The expression results from the particular form of the diﬀerential equation to be solved (second derivative and periodic potential) and is justiﬁed using the Floquet theorem. Expression (8.21) of ψk (x), established in the case of weak coupling between neighboring atomic sites, can be interpreted as a Linear Combination of Atomic Orbitals (LCAO), since the φ0 ’s then do not diﬀer much from the atomic solutions ϕ. This model is also called “tight binding” as it privileges the atomic states. When the atomic states of two neighboring sites are taken as orthogonal, this is the so-called “Hückel model.” The ψk (x) functions are delocalized all over the crystal. The extreme energy states, at the edges of the energy band, correspond to particular Bloch functions : the minimum energy state ε0 − 2A is associated with ψ0 (x) = +∞ n=−∞ φ0 (x−nd), the addition in phase of each localized wave function : this is a situation comparable to the bonding symmetricalstate of the H+ 2 ion ; the n state of energy ε0 + 2A is associated with ψπ/d (x) = +∞ n=−∞ (−1) φ0 (x− nd), alternate addition of each localized wave function and is analogous to the antibonding state of the H+ 2 ion.

184

Chapter 8. Elements of Bands Theory and Crystal Conductivity

One can notice that the band width 4A obtained for an inﬁnite chain is only twice the value of the energy splitting for a pair of coupled sites. The reason is that, in the present model, a given site only interacts with its ﬁrst neighbors and not with the more distant sites Note : the dispersion relation (8.20) and the wave function (8.24), that we obtained under very speciﬁc hypotheses, can be generalized in the case of three-dimensional periodic solids, as can be found in basic Solid State Physics courses, like the one taught at Ecole Polytechnique.

8.3 8.3.1

The Electron States in a Crystal Wave Packet of Bloch Waves

Bloch waves of the (8.21) or (8.24) type are delocalized and thus cannot describe the amplitude of probability of presence for an electron, as these waves cannot be normalized : although a localized function φ0 (x) can be normed, ψk (x) spreads to inﬁnity. To describe an electron behavior, one has to build a packet of Bloch waves

+∞

Ψ(x, t) = −∞

g(k)eikx uk (x)e−iε(k)t/ dk

(8.26)

centered around a wave vector k0 , of extent ∆k small with respect to π/d. From the Heisenberg uncertainty principle, the extent in x of the packet will be such that ∆x · ∆k ≥ 1, i.e., ∆x d : the Bloch wave packet is spreading over a large number of elementary cells.

8.3.2

Resistance ; Mean Free Path

It can be shown that an electron described by a wave packet (8.26), in motion in a perfect inﬁnite crystal, keeps its velocity value for ever and thus is not diﬀused by the periodic lattice ; in fact the Bloch wave solution already includes all these coherent diﬀusions. Then there is absolutely no resistance to the electron motion or, in other words, the conductivity is inﬁnite. modiﬁes the electron energy levels ; the proA static applied electric ﬁeld E blem is complex, one should also account for the eﬀect of the lattice ions and interact with the electron. which are also subjected to the action of E Experimentally one notices in a macroscopic crystal that the less its defects,

The Electron States in a Crystal

185

like dislocations or impurities, the higher its conductivity. An inﬁnite perfect three-dimensional crystal would have an inﬁnite conductivity, a result predicted by Felix Bloch as early as 1928. (These simple considerations do not apply to the recently elaborated, low-dimensional, systems.) The real ﬁnite conductivity of crystals arises from the discrepancies to the perfect inﬁnite lattice : thermal motion of the ions, defects, existence of the surface. In fact in copper at low temperature one deduces from the conductivity value σ = ne2 τ /m the mean time τ between two collisions and, introducing the Fermi velocity vF (see §7.1.1), one obtains a mean free path = vF τ of the order of 40 nm, that is, about 100 interatomic distances. This does show that the lattice ions are not responsible for the diﬀusion of conduction electrons.

8.3.3

Finite Chain, Density of States, Eﬀective Mass

V (x) 0

d

2d

Nd

(N + 1)d x

Fig. 8.5 : Potential for a N -ions chain. A real crystal is spreading over a macroscopic distance L = N d (Fig. 8.5) (the distance L is measured with an uncertainty much larger than the interatomic distance d ; one can thus always assume that L contains an integer number N of interatomic distances). Obviously, if only the coupling to the nearest neighbors is considered, for most sites there is no change with respect to the case of the inﬁnite crystal. Thus the eigenstates of the hamiltonian for the ﬁnite crystal : 2 ˆ crystal = pˆ + V (x − xn ) h 2m n=1 N

(8.27)

should not diﬀer much from those of the hamiltonian (8.9) for the inﬁnite crystal : we are going to test the Bloch function obtained for the inﬁnite crystal as a solution for the ﬁnite case. On the other hand, for a macroscopic dimension L the exact limit conditions do not matter much. It is possible, as in the case of the free electron model, to close the crystal on itself by superposing the sites 0 and N d and to express

186

Chapter 8. Elements of Bands Theory and Crystal Conductivity

the periodic limit conditions of Born-Von Kármán : ψ(x + L) = ψ(x) ψk (x + L) = e

ik(x+L)

(8.28) uk (x + L) = e

ikL

ψk (x)

(8.29)

as L contains an integer number of atoms, i.e. an integer number of atomic periods d, and uk (x) is periodic. One thus ﬁnds the same quantization condition as for a one-dimensional free particle : k = p2π/L, where p is an integer, associated with an one-dimension orbital density of states L/2π ; the corresponding three-dimensional density of states is Ω/(2π)3 , where Ω is the crystal volume. These densities are doubled when one accounts for the electron spin. The solutions of (8.27) are then ⎧ kd ⎪ ⎨ε = ε0 − 2A cos N 1 ψk (x) = √N n=1 exp iknd · φ0 (x − nd) ⎪ ⎩ with k = p · 2π p integer L

(8.30)

When comparing to the results (8.20) and (8.21) for the inﬁnite crystal, one ε is now quanﬁrst notices that ψk (x) as given by (8.30) is normed. Moreover π π tized (Fig. 8.6) : to describe the ﬁrst Brillouin zone − ≤ k < all the d d N N diﬀerent states are obtained for p ranging between − and + − 1 (or 2 2 N −1 N −1 between − and − 1 if N is odd, but because of the order of 2 2 magnitude of N , ratio of the macroscopic distance L to the atomic one d, this does not matter much !). On this interval there are thus N possible states, i.e., as many as the number of sites. Half of the states correspond to waves propagating toward increasing x’s (k > 0), the other ones to waves propagating toward decreasing x’s. Note that if the states are uniformly spread in k, their distribution is not uniform in energy. One deﬁnes the energy density of states D(ε) such that the number of accessible states dn, of energies between ε and ε + dε, is equal to dn = D(ε)dε

(8.31)

One will note that, at one dimension, to a given value of ε correspond two opposite values of kx , associated with distinct quantum states. By observing Fig. 8.6, it can be seen that D(ε) takes very large values at the vicinity of the band edges (ε close to ε0 ± 2A) ; in fact at one dimension D(ε) has a speciﬁc behavior [D(ε) diverges at the band edges, but the integral ε D(ε )dε (8.32) εmin

The Electron States in a Crystal

187

ε(k) ε +2A

ε0 − 2 A −π/d

2π 2π = Nd L

0

π/d

k

Fig. 8.6 : Allowed states in the energy band, for a N electrons system.

remains ﬁnite]. When this model is extended to a three-dimensional system, one ﬁnds again a “quasicontinuous” set of very close accessible states, uniformly located in k (but not in energy), in number equal to that of the initial orbitals (it was already the case for two atoms for which we obtained the bonding and antibonding states). V x 2s

1s

2s band

1s band

Fig. 8.7: In a N atom crystal, to each atomic state corresponds a band that can accommodate 2N times more electrons than the initial state. In a more realistic case, the considered atom has several energy levels (1s, 2s, 2p, ...). Each energy level gives rise to a band, wider if the atomic state lies higher in energy (Fig. 8.7). The widths of the bands originating from the deep atomic states are extremely small, one can consider that these atomic levels do not couple. On the other hand, the width of the band containing the electrons involved in the chemical bond is always of the order of several electron-volts.

188

Chapter 8. Elements of Bands Theory and Crystal Conductivity

Again each band contains N times more orbital states than the initial atomic state, that is, 2N times more when the electron spin is taken into account. The density of states in k is uniform. One deﬁnes an energy density of states D(ε) such that Å dn = D(ε)dε = 2

L 2π

ã3

d3k at three dimensions

(8.33)

The area under the curve representing the density of states of each band is equal to 2N times the number of electrons acceptable in the initial atomic orbital : 2N for s states, 6N for p states, etc. (Fig. 8.8). The allowed bands are generally separated by forbidden bands, except when there is an overlap between the higher energy bands, which are broadened through the coupling (3s and 3p on the Fig. 8.8). 1s D(ε)

2p 2s 3s, 3p

ε Fig. 8.8 : Density of states of a crystal. It can be shown that in the vicinity of a band extremum, the tangent of the D(ε) curve is vertical. In particular, at the bottom of a band the density of state is rather similar to the one of a free electron in a box [see (6.64)], for √ Ω which we obtained D(ε) = 2 3 (2m)3/2 ε, m being the free electron mass. 2π Here again one will deﬁne the eﬀective mass in a given band, this time from the curvature of D(ε) [this mass is the same as the one in § 8.2.3, introduced from the dispersion relation (8.20)]. Indeed, setting D(ε)

√ Ω (2m∗ )3/2 ε − ε0 2π 2 3

(8.34)

in the vicinity of ε0 at the bottom of the band is equivalent to taking for dispersion relation ε − ε0

2 k 2 2m∗

,

i.e.,

2 d2 ε = m∗ dk 2

(8.35)

Statistical Physics of Solids

189

The eﬀective mass is linked to the band width : from the dispersion relation (8.20) for instance, one obtains around k = 0 2Ad2 1 = ∗ m 2

(8.36)

One can understand that the electrons of the deep bands are “very heavy,” and thus not really mobile inside the lattice, that is, they are practically localized. Note : in non crystalline (or amorphous) solids, like amorphous silicon, the coupling between the atomic states also produces energy bands.

8.4

Statistical Physics of Solids

Once the one-electron quantum states determined, one has to locate the N electrons in these states, in the framework of the independent electrons model.

8.4.1

Filling of the Levels

At zero temperature, the Pauli principle dictates one to ﬁll the levels from the lowest energy one, putting two electrons, with diﬀerent spin states, on each orbital state until the N electrons are exhausted. We now take two examples, that of lithium (Z = 3), and that of carbon (Z = 6) crystallized into the diamond form. In lithium, each atom brings 3 electrons, for the whole solid there are 3N electrons to be located. The 1s band is ﬁlled by 2N electrons. The complementary N electrons are ﬁlling the 2s band up to its half, until the energy ε = εF , the Fermi energy, which lies 4.7 eV above the bottom of this band. Above εF , any state is empty at T = 0 K. When a small amount of energy is brought to the solid, for example using a voltage source of a few volts, the electrons in the 2s band in the vicinity of εF can occupy states immediately above εF , whereas 1s electrons cannot react, having no empty states in their neighborhood. This energy modiﬁcation of electrons in the 2s band produces a global macroscopic reaction, expressed by an electric current, to the electric ﬁeld : lithium is a metal or a conductor, the 2s band is its conduction band. In diamond, the initial electronic states are not the 2s and 2p states of the atomic carbon but their sp3 tetrahedral hybrids (as in CH4 ) which separate into bonding hybrids with 4 electrons per atom (including spin) and antibon-

190

Chapter 8. Elements of Bands Theory and Crystal Conductivity

ding hybrids also with 4 electrons. In the crystal, the bonding states constitute a 4N -electron band, the same is true for the antibonding states. The 4N electrons of the N atoms exactly ﬁll the bonding band, which contains the electrons participating in the chemical bond, whence the name of valence band. The band immediately above, of antibonding type, is the conduction band. It is separated from the valence band by the band gap energy, noted Eg : Eg = 5.4 eV for diamond and at T = 0 K the conduction band is empty, the valence band having exactly accommodated all the available electrons. Under the eﬀect of an applied voltage of a few volts, the valence electrons cannot be excited through the forbidden band, there is no macroscopic reaction, the diamond is an insulator. Like carbon, silicon and germanium also belong to column IV of the periodic table, they have the same crystalline structure, they are also insulators. However these latter atoms are heavier, their crystal unit cells are larger and the characteristic energies smaller : thus Eg = 1.1 eV in silicon and 0.75 eV in germanium. Remember that a full band has no conduction, the application of an electric ﬁeld producing no change of the macroscopic occupation of the levels. This is equally valid for the 1s band of the lithium metal and at zero temperature for the valence band of insulators like Si or Ge.

8.4.2

Variation of Metal Resistance versus Temperature

We now are able to justify the methods and results on metals of the previous chapter, by reference to the case of lithium discussed above : their bands consist in one or several full bands, inactive under the application of an electric ﬁeld, and a conduction band partly ﬁlled at zero temperature. The density of states at the bottom of this latter band is analogous to that of an electron in a box, under the condition that the free electron mass is replaced by the eﬀective mass at the bottom of the conduction band. Indeed we already anticipated this result when we analyzed the speciﬁc heats of copper and silver (7.2.2). The electrical conductivity σ is related to the linear eﬀect of an applied ﬁeld E on the conduction electrons. An elementary argument allows one to establish its expression σ = ne2 τ /m∗ , where n is the electron concentration in the conduction band, e the electron charge, τ the mean time between two collisions and m∗ the eﬀective mass. A more rigorous analysis, in the framework of the transport theory, justiﬁes this formula which is far from being obvious : the only excitable electrons are in the vicinity of the Fermi level, their velocity in the absence of electric ﬁeld is vF . The phenomenological τ term takes into account the various mechanisms

Statistical Physics of Solids

191

which limit the conductivity : in fact the probabilities of these diﬀerent independent mechanisms are additive, so that 1 1 1 = + + ··· τ τvibr τimp

(8.37)

1/τvibr corresponds to the probability of diﬀusion by the lattice vibrations, which increases almost linearly with the temperature T , 1/τimp to the diffusion by neutral or ionized impurities, a mechanism present even at zero temperature. Then 1 m∗ =ρ= 2 σ ne

Å

1 τvibr

+

1 τimp

ã + ···

(8.38)

ρ = ρvibr (T ) + ρimp + · · · Thus the resistivity terms arising from the various mechanisms add up, this is the empirical Matthiessen law, stated in the middle of the 19th century. Consequently, the resistance of a metal increases with temperature according to a law practically linear in T , the dominating mechanism being the electron’s diﬀusion on the lattice vibrations. This justiﬁes the use of the platinum resistance thermometer as a secondary temperature standard between −183◦ C and +630◦ C.

8.4.3

Variation of Insulator’s Conductivity Versus Temperature ; Semiconductors

a) Concentrations of Mobile Carriers in Equilibrium at Temperature T We have just seen that at zero temperature an electric ﬁeld has no eﬀect on an insulator, since its valence band is totally full and its conduction band empty. However there are processes to excite such systems. The two main ones are : – an optical excitation : a photon of energy hν > Eg will promote an electron from the valence band into the conduction band. This is the absorption phenomenon, partly responsible for the color of these solids. It creates a nonequilibrium thermal situation. – a thermal excitation : the thermal motion is also at the origin of transitions from the valence band to the conduction band. We are now going to detail this latter phenomenon, which produces a thermal equilibrium at temperature T .

192

Chapter 8. Elements of Bands Theory and Crystal Conductivity

When an electron is excited from the valence band, this band then makes a transition from the state (full band) to the state (full band minus one electron), which allows one to deﬁne a quasi-particle, the “hole,” which is the lack of an electron. We already noted in chapter 7.1.1 the symmetry, due to the shape of the Fermi distribution, between the occupation factor f (ε) of energy states above µ, and the probability 1 − f (ε) of the states below µ for being empty. The following calculation will show the similarity between the electron or hole role. The electrons inside the insulator follow the Fermi-Dirac statistics and at zero temperature the conduction band is empty and the valence band full. This means that εF lies between εv , the top of the valence band and εc , the minimum of the conduction band. At weak temperature T (kB T Eg ) one expects that the chemical potential µ still lies inside the energy band gap. We are going to estimate n(T ), the number of electrons present at T in the conduction band and p(T ), the number of electrons missing (or holes present) in the valence band, and to deduce the energy location µ(T ) of the chemical potential. These very few carriers will be able to react to an applied electric : indeed in the conduction band the electrons are very scarce at usual ﬁeld E temperatures and they are able to ﬁnd vacant states in their close neighbo ; in the same way the presence of several holes in rhood when excited by E the valence band permits the valence electrons to react to E. The volume concentration n(T ) of electrons present in the conduction band at temperature T in the solid of volume Ω is such that εcmax εcmax 1 dε (8.39) D(ε)f (ε)dε = D(ε) Ωn(T ) = exp β(ε − µ) + 1 εc εc where εc is the energy of the conduction band minimum, εcmax the energy of its maximum. One assumes, and it will be veriﬁed at the end of the calculation, that µ is distant of this conduction band of several kB T , which leads to β(εc − µ) 1, so that one can approximate f (ε) by a Boltzmann factor. Then Ωn(T )

εcmax

D(ε)e−β(ε−µ) dε

(8.40)

εc

Because of the very fast exponential variation, the integrand has a signiﬁcant value only if ε − εc does not exceeds a few kB T . Consequently, one can extend the upper limit of the integral to +∞, as the added contributions are absolutely negligible. Moreover, only the expression of D(ε) at the bottom of the conduction band really matters : we saw in (8.34) that an eﬀective mass, here mc , can be introduced, which expresses the band curvature in the vicinity of

Statistical Physics of Solids

193

ε = εc . Whence

∞

√ 1 (2mc )3/2 ε − εc e−β(ε−µ) dε 2 3 2π εc ∞ √ −x 1 −β(εc −µ) 3/2 e (2mc kT ) xe dx = 2π 2 3 0 ã ã Å Å 2πmc kB T 3/2 εc − µ =2 exp − h2 kB T

n(T )

(8.41)

(8.42)

In the same way one calculates the hole concentration p(T ) in the valence band, the absence of electrons at temperature T being proportional to the factor 1 − f (ε) : εv Ωp(T ) = D(ε)[1 − f (ε)]dε (8.43) εvmin

One assumes that µ is at a distance of several kB T from εv , the maximum of the valence band, of minimum εvmin . The neighborhood of εv is the only region that will intervene in the integral calculation because of the fast decrease of 1 − f (ε) e−β(µ−ε) . In this region D(ε) is parabolic and is approximated by D(ε)

√ Ω (2mv )3/2 εv − ε 2 3 2π

(8.44)

where mv > 0, the “hole eﬀective mass,” expresses the curvature of D(ε) in the neighborhood of the valence band maximum. One gets : Å

2πmv kB T p(T ) = 2 h2

ã3/2

ã Å µ − εv exp − kB T

(8.45)

In the calculations between formulae (8.40) and (8.45) we did not make any assumption about the way n(T ) and p(T ) are produced. These expressions are general, as is the product of (8.42) by (8.45) : Å n(T ) · p(T ) = 4

2πkB T h2

ã3

ã Å Eg (mc mv )3/2 exp − kB T

(8.46)

with Eg = εc − εv . b) Pure (or Intrinsic) Semiconductor In the case of a thermal excitation, for each electron excited into the conduction band a hole is left in the valence band. One deduces from (8.46) : Å n(T ) = p(T ) = ni (T ) = 2

2πkB T h2

ã3/2

ã Å Eg (mc mv )3/4 exp − 2kB T

(8.47)

194

Chapter 8. Elements of Bands Theory and Crystal Conductivity

For this intrinsic mechanism, the variation of ni (T ) versus temperature is dominated by the exp(−Eg /2kB T ) term. In fact, here the thermal mechanism creates two charge carriers with a threshold energy Eg . The very fast variation of the exponential is responsible for the distinction between insulators and semiconductors. Indeed at room temperature T = 300 1 eV, let us compare the diamond K, which corresponds to approximately 40 situation where Eg = 5.2 eV [exp(−Eg /2kB T ) = exp(−104) = 6.8 × 10−46 ] and that of silicon, where Eg = 1.1 eV [exp(−22) = 2.8 × 10−10 ]. In 1 cm3 of diamond, there is no electron in the conduction band, whereas ni (T ) is equal to 1.6 × 1016 m−3 at 300 K in silicon. A semiconductor is an insulator with a “not too wide” band gap, which allows the excitation of some electrons into the conduction band at room temperature. The most commonly used semiconductors, apart from silicon, are germanium and gallium arsenide GaAs (Eg = 1.52 eV). The chemical potential of an intrinsic semiconductor is determined by substituting expression (8.47) of ni (T ) into (8.41). One obtains : µ=

3 mv εc + εv + kB T ln 2 4 mc

(8.48)

The Fermi level εF , which is the chemical potential at T = 0 K, lies in the middle of the band gap and µ does not move much apart from this location when T increases (kB T Eg ). Consequently, the hypotheses that µ lies very far from the band edges as compared to kB T , made when calculating n(T ) and p(T ), are valid. Under the eﬀect of an external electric ﬁeld, both electrons and holes participate in the electrical current : indeed the current includes contributions from both the conduction and the valence bands. As will be studied in Solid State Physics courses, the holes behave like positive charges and the total electrical conductivity in the semiconductor is the sum of the valence band and conduction band currents : σ = σn + σp ã Å τn τp = ne2 + pe2 mc mv

(8.49)

Here τn (respectively τp ) is the electron (respectively hole) collision time. The temperature variation of σ in an intrinsic semiconductor is dominated by the variation of exp(−Eg /2kT ), the other factors varying like powers of T . Consequently, σ increases very fast with T , thus its reciprocal, the resistivity and the measured resistance of a semiconductor decrease with temperature because the number of carriers increases.

Statistical Physics of Solids

195

c) Doped Semiconductor The density of intrinsic carriers at room temperature is very weak and practically impossible to attain : for instance in silicon about 1 electron out of 1012 is in the conduction band due to thermal excitation. To achieve the intrinsic concentration the studied sample should contain neither impurities nor defects ! It is more realistic to control the conductivity by selected impurities, the dopants. Thus phosphorus atoms can be introduced into a silicon crystal : phosphorus is an element of column V of the periodic table, which substitutes for silicon atoms in some lattice sites. Each phosphorus atom brings one more electron than required by the tetrahedral bonds of Si. At low temperature this electron remains localized near the phosphorus nucleus, on a level located 44 meV below the bottom of the conduction band. When the temperature increases, the electron is ionized into the conduction band. Phosphorus is a donor. The concentration of conduction electrons will be determined by the concentration ND in P atoms, of typical value ND ∼ 1022 m−3 , chosen much larger than the intrinsic and the residual impurity concentrations, but much smaller than the Si atoms concentration. In such a material, because of (8.46), the hole concentration is very small with respect to ni , the current is carried by the electrons, the material is called of n-type. In a symmetrical way, the substitution in the silicon lattice of a Si atom by a boron atom creates an acceptor state : boron, an element of column III of the periodic table, only has three peripheral electrons and thus must capture one extra electron in the valence band to satisfy the four tetrahedral bonds of the crystal. This is equivalent to creating a hole, which, at low temperature, is localized 46 meV above the silicon valence band and at higher temperature is delocalized inside the valence band. In boron-doped silicon, the conductivity arises from the holes, the semiconductor is of p-type. In Solid State Physics courses, the statistics of doped semiconductors is studied in more detail : one ﬁnds in particular that the chemical potential µ moves versus temperature inside the band gap (in a n-type material, at low temperature it is close εc , in the vicinity of the donors levels and returns to the middle of the band gap, like in an intrinsic material, at very high temperature). The semiconductors constituting the electronic compounds are generally doped.

196

8.5

Chapter 8. Elements of Bands Theory and Crystal Conductivity

Examples of Semiconductor Devices

Here we only give a ﬂavor of the variety of conﬁgurations and uses. In specialized Solid State Physics courses, one will describe in more details the semiconductors properties. The aim here is to show that, just using the basic elements of the present chapter, one can understand the principle of operation of three widespread devices.

8.5.1

The Photocopier : Photoconductivity Properties

Illuminating a semiconductor by light of energy larger than Eg promotes an electron into the conduction band and leaves a hole in the valence band. The conductivity increase is ∆σ, proportional to the light ﬂux. In the xerography process, the photocopier drum is made of selenium which, under the illumination eﬀect, shifts from a very insulator state to a conducting state. This property is utilized to ﬁx the pigments (the toner black powder) which are next transferred to the paper and then cooked in an oven to be stabilized on the paper sheet.

8.5.2

The Solar Cell : an Illuminated p − n Junction

When, in the same solid, a p region is placed next to a n region, the electrons of the n region diﬀuse toward the holes of the p region : some of these mobile charges neutralize and a so-called space charge region becomes emptied of mobile carriers. The dopants still remain in the region, their charged nuclei create an electric ﬁeld which prevents any additional diﬀusion. This eﬀect can also be interpreted from the requirement of equal chemical potentials in the n and p region at equilibrium : in the neighborhood (a few µm) of the i . If region of doping change, there prevails an intense internal electric ﬁeld E the sun is illuminating this region, the excited electron and hole, of opposite i . An external use circuit will take advantage of charges, are separated by E the current produced by this light excitation. We will reconsider these devices, called “solar cells,” at the end of next chapter when we will deal with thermal radiation.

Examples of Semiconductor Devices

8.5.3

197

Compact Disk Readers : Semiconductor Quantum Wells

It was seen in § 8.3.3 that starting from the energy levels of the constituting atoms one gets the energy bands of a solid. Their location depends of the initial atomic levels, thus of the chemical nature of the material. For almost thirty years it has been possible, in particular using Molecular Beam Epitaxy (MBE), to sequentially crystallize diﬀerent semiconductors, which have the same crystal lattice but diﬀerent chemical composition : a common example is GaAs and Ga1−x Alx As. In such semiconductor sandwiches the bottom of the conduction band (or the top of the valence band ) does not lie at the same energy in GaAs and in Ga1−x Alx As. One thus tailors quantum wells, to fashion. Their width usually ranges from 5 to 20 nm, their energy depth is of a few hundreds of meV, and the energies of their localized levels is determined by their “geometry,” as taught in Quantum Mechanics courses. These quantum wells are much used in optoelectronics, for example as diodes or lasers emitters, in the telecommunication networks by optical ﬁbers, as readers of compact disks, of code-bars in supermarkets, etc.

Summary of Chapter 8 One has ﬁrst to study the quantum states of a periodic solid. By analogy with the problem of the double potential well of the Quantum Mechanics course, one considers an electron on an inﬁnite periodic chain, in a model of independent electrons, of the Linear Combination of Atomic Orbitals type. The hamiltonian eigenstates are Bloch states, the eigenenergies are regrouped into allowed bands separated by forbidden bands (or band gaps). The important parameter is the coupling term between neighboring sites. In a perfect crystal, the velocity of an electron would be maintained forever. The ﬁnite macroscopic dimension of a crystal allows one to deﬁne a density of states and to show that the number of orbital states is the product of the number of cells in the crystal by the number of orbital states corresponding to each initial atomic state. Filling these states together with considering the spin properties, allows one to distinguish between insulators and metals. We calculated the number of mobile carriers of an insulator versus temperature. The resistance of an insulator exponentially decreases when T increases, the resistance of a metal increases almost linearly with T . A semiconductor is an insulator with a “rather small” band gap energy.

199

Chapter 9

Bosons : Helium 4, Photons, Thermal Radiation We learnt in chapter 5 that bosons are particles of integer or null spin, with a symmetrical wave function, i.e. which is unmodiﬁed in the exchange of two particles. Bosons can be found in arbitrary number on a one-particle state of given energy εk . We are going now to discuss in more detail the properties of systems following the Bose-Einstein statistics. We will begin with the case of material particles and take the example of 42 He, the helium isotope made up of two protons and two neutrons. The Bose-Einstein condensation occurring in this case is related to the helium property of superﬂuidity. Then we will study the statistics in equilibrium of photons, relativistic particles of zero mass. The photon statistics has an essential importance for our life on the earth through the greenhouse eﬀect. Moreover, it played a major conceptual part : indeed it is in order to interpret the experimental thermal radiation behavior that Max Planck (1858-1947) produced the quanta hypothesis (1901). Finally we will take a macroscopic point of view and stress the importance of thermal radiation phenomena in our daily experience.

9.1

Material Particles

The ability of bosons of given non-zero mass to gather on the same level under a low enough temperature yields to a phase transition, the Bose-Einstein condensation.

201

202

9.1.1

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Thermodynamics of the Boson Gas

Let us recall the average occupation number nk [Eq.(6.26)] by bosons of 1 : chemical potential µ, of the energy level εk at temperature T = kB β 1

nk =

eβ(εk −µ)

−1

(9.1)

As nk is positive or zero, the denominator exponent should be positive for any εk , which implies that the chemical potential µ must be smaller than any εk , including ε1 , the energy of the fundamental state. At ﬁxed T and µ, nk decreases when εk increases. The µ value is ﬁxed by the condition

nk = N

(9.2)

k

where N is the total particles number of the considered physical system. From nk one can calculate all the thermodynamical parameters and in particular the average energy : U=

εk nk

(9.3)

k

the grand potential A and its partial derivatives S and P A = −kB T ln(1 + nk )

(9.4)

k

dA = −P dΩ − SdT − N dµ

(9.5)

In the large volume limit, (9.1) is replaced by the Bose-Einstein distribution, a continuous function given by f (ε) =

1 eβ(ε−µ) − 1

(9.6)

The physical parameters are calculated in this case using the density of states D(ε), then expressing the constraint replacing (9.2) on the particles number and verifying that µ < ε1 . Here (§ 9.1) one is concerned by free bosons, with no p2 (apart from mutual interaction, therefore their energy is reduced to ε = 2m the “box” potential producing the quantization, leading to the expression of D(ε)), so that ε1 = 0 and µ < 0. Then for particles with a spin s, in the three-dimensional space, one gets,

Material Particles

203

Å ã √ 2s + 1 2m 3/2 taking D(ε) = CΩ ε, where C = [cf. Eq. (6.64)], 4π 2 2 √ ∞ ∞ εdε N= f (ε)D(ε)dε = CΩ β(ε−µ) e −1 0 0 ∞ ∞ 3/2 ε dε U= εf (ε)D(ε)dε = CΩ β(ε−µ) − 1 e 0 0 ∞ A = −kB T ln[1 + f (ε)]D(ε)dε 0 ∞ √ ε ln[1 − e−β(ε−µ) ]dε A = CΩkB T 2 = − CΩ 3

0 ∞

0

ε3/2 dε −1

eβ(ε−µ)

(9.7) (9.8)

(9.9) (9.10)

Integrating (9.9) by parts, one ﬁnds (9.10) which is equal to (9.8) to the factor 2 − , so that 3 2 A = − U = −P Ω (9.11) 3 This is the state equation of the bosons gas, that is, 2U (9.12) 3Ω which takes the same form as for the fermions (7.13), although the physical properties are totally diﬀerent. P =

9.1.2

Bose-Einstein Condensation

The above expressions are only valid when the condition µ < 0 is fulﬁlled (see above). Now when T decreases at constant density N/Ω, since β increases the diﬀerences (ε − µ) must decrease in order that N/Ω keeps its value (9.7). This means that µ gets closer to zero. Now consider the limit situation µ = 0. Relation (9.7) then becomes ∞ √ εdε N = CΩ βε − 1 e 0 √ Å ã CΩ ∞ xdx CΩ 1 N = 3/2 = 3/2 I x−1 e 2 0 βB βB

(9.13) (9.14)

i.e., using the mathematical tables of the present book, N=

CΩ 3/2

βB

· 2, 315

(9.15)

204

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Relation (9.15) deﬁnes a speciﬁc temperature TB = 1/kB βB which is only related to the density n = N/Ω : ã2/3 Å N 1 1 kB CΩ 2, 315 n2/3 6, 632 = 2mkB (2s + 1)2/3

TB =

In (9.17) C was replaced by its expression C =

2s + 1 4π 2

spin of the 42 He particles is zero.

(9.16) (9.17) Å

2m 2

ã3/2 , in which the

The temperature TB , or Bose temperature, would correspond in the above calculation to µ = 0, whereas the situation of § 9.1.1 is related to µ < 0. Since the integrals in (9.7) and (9.13) are both equal to N/CΩ, to maintain this value constant it is necessary that β (corresponding to µ < 0) be smaller than βB (associated to µ = 0). Consequently, (9.7) is only valid for T > TB . When lowering the temperature to TB , µ tends to zero, thus getting closer to the fundamental state of energy ε1 = 0. For T < TB the analysis has to be reconsidered : the population of level ε1 has to be studied separately, as it would diverge if one were using expression (9.1) without caution ! Now the population N1 of ε1 is at most equal to N . Expression (9.1) remains valid but µ cannot be strictly zero. Assume that N1 is a macroscopic number of the order of N : N1 =

1 1 = −βµ e −1 eβ(ε1 −µ) − 1

(9.18)

The corresponding value of µ is given by : 1 N1 kB T µ− N1

e−βµ = 1 +

(9.19)

µ is very close to zero because it is the ratio of a microscopic energy to a macroscopic number of particles. The ﬁrst excited level ε2 has the energy 2 ∝ Ω−2/3 , whereas the chemical potential is in Ω−1 , so that µ is very 2mL2 small as compared to ε2 . The N − N1 particles which are not located in ε1 are distributed among the excited states, for which one can take µ = 0 without introducing a large error,

Material Particles

205

and adapt the condition (9.7) on the particle number ∞ √ εdε N − N1 = CΩ β − 1 e ε 2∞ √ εdε CΩ βε − 1 e ε1 =0

(9.20) (9.21)

Indeed shifting the limit of the integral from ε2 to ε1 is adding a relative contribution to (9.20) of the order of (ε2 /kB T )1/2 . Since the volume Ω is macroscopic, this quantity is very small, ε2 being much less than kB T . One recognizes in (9.21) the integral (9.13) under the condition of replacing βB by β. Whence : N − N1 T 3/2 ñ Å ã ô T 3/2 N1 = N 1 − TB

N

3/2 TB

=

(9.22) (9.23)

Thus, when lowering the temperature, diﬀerent regimes take place : – as long as T > TB , the chemical potential is negative and the particles are distributed among all the microscopic states of the system according to the Bose-Einstein distribution (9.6) ; – when T = TB , µ almost vanishes (it is very small but negative) ; – when T < TB , the particles number N1 on the fundamental level becomes macroscopic, this is the “Bose condensation.” One thus realizes a macroscopic quantum state. The distribution of particles on the excited levels corresponds to µ 0. When T continues to decrease, N1 increases and tends to N when T tends to zero ; µ is practically zero. Going through TB corresponds to a phase transition : for T < TB the Bose gas is degenerate, which means that a macroscopic particles number is located in the same quantum state. This corresponds for 42 He to the superﬂuid state. The phenomenon of helium superﬂuidity is very complex and the interactions between atoms, neglected up to now, play a major part in this phase. The above calculation predicts a transition to the superﬂuid state at 3.2 K in standard conditions (4 He speciﬁc mass =140 kg/m3 under the atmospheric pressure) whereas the transition occurs at Tλ = 2.17 K. This is a second-order phase transition (the free energy is then ﬁnite, continuous and derivable ; the speciﬁc heat is discontinuous). The superconductivity of some solids also corresponds to a Bose condensation (2003 Nobel prize of A. Abrikosov and V. Ginzburg). It was possible to observe the atoms’ Bose condensation by trapping and cooling atomic beams at temperatures of the order of a mK in “optical molasses” (1997 Nobel prices of S. Chu, C. Cohen-Tannoudji and W.D. Phillips).

206

9.2

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Bose-Einstein Distribution of Photons

Our daily experience teaches us that a heated body radiates, as is the case for a radiator heating a room or an incandescent lamp illuminating us through its visible radiation. Understanding these mechanisms, a challenge at the end of the 19th century, was at the origin of M. Planck’s quanta hypothesis (1901). For more clarity we will not follow the historical approach but rather use the tools provided by this course. Once the results obtained, we will return to the problems Planck was facing.

9.2.1

Description of the Thermal Radiation ; the Photons

Consider a cavity of volume Ω heated at temperature T (for example an oven) and previously evacuated. The material walls of this cavity are made of atoms, the electrons of which are promoted into excited states by the energy from the heat source. During their deexcitation, these electrons radiate energy, under the form of an electromagnetic ﬁeld which propagates inside the cavity, is absorbed by other electrons associated to other atoms, and so forth, and this process leads to a thermal equilibrium between the walls and the radiation enclosed inside the cavity. If the cavity walls are perfect reﬂectors, classical electromagnetism teaches us is zero inside the perfect conductor and normal that the wave electrical ﬁeld E tangential is contito the wall on the vacuum side (due to the condition that E nuous), and that the corresponding magnetic ﬁeld B must be tangential. In order to produce the resonance of this cavity and to establish stationary waves, the cavity dimension L should contain an integer number of half-wavelengths λ. The wave vector k = 2π/λ and L are then linked. k) exp i(k ·r −ωt) An electromagnetic wave, given in complex notation by E(ω, is characterized by its wave vector k and its frequency ω ; it can have two independent polarization states (either linear polarizations along two perpendicular directions, or left or right circular polarizations) on which the direction of the electric ﬁeld in the plane normal to k is projected. Finally its intensity k)|2 . is proportional to |E(ω, Since the interpretation of the photoelectric eﬀect by A. Einstein in 1905, it has been known that the light can be described both in terms of waves and of particles, which are the photons. The quantum description of the electromagnetic ﬁeld is taught in advanced Quantum Mechanics courses. Here we only specify that the photon is a relativistic particle of zero mass, so that

Bose-Einstein Distribution of Photons

207

the general relativistic relation between the energy ε and the momentum p, ε = p2 c2 + m20 c4 , where c is the light velocity in vacuum, reduces in this case to ε = pc. Although the photon spin is equal to 1, due to its zero mass it only has two distinct spin states. The set of data consisting in p and the photon spin value deﬁnes a mode. The wave parameters k and ω and the photon parameters are related through the Planck constant, as summarized in this table : electromagnetic wave wave vector k frequency ω k, ω)|2 intensity |E( polarization (2 states)

particle : photon p = k momentum ε = ω energy number of photons spin (2 values)

Let us return to the cavity in thermal equilibrium with the radiation it contains : its dimensions, in Ω1/3 , are very large with respect to the wavelength of the considered radiation. The exact surface conditions do not matter much : as in the particles case (see § 6.4.1b) ), rather than using the stationary wave conditions, we prefer the periodic limit conditions (“Born-Von Kármán” conditions) : in a thought experiment we close the system on itself, which quantizes k and consequently, p and the energy. Assuming the cavity is a box of dimensions Lx , Ly , Lz , one obtains : ã Å k = (kx , ky , kz ) = 2π nx , ny , nz (9.24) Lx Ly Lz where nx , ny , nz are positive or negative integers, or zero. Then

Å p = h

nx ny nz , , Lx Ly Lz

ε = pc = ω = hν

9.2.2

ã (9.25) (9.26)

Statistics of Photons, Bosons in Non-Conserved Number

In the photon emission and absorption processes of the wall atoms, the number of photons varies, and there are more photons when the walls are hotter : in this system the constraint on the conservation of particles number, that existed in § 9.1, is lifted. This constraint was deﬁning the chemical potential µ as a Lagrange multiplier, which no longer appears in the probability law for photons, i.e., the chemical potential is zero for photons.

208

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

Let us consider the special case where only a single photon mode is possible, i.e., a single value of p, of the polarization and of ε in the cavity at thermal equilibrium at temperature T . Any number of photons is possible, thus the partition function z(ε) is equal to z(ε) =

∞

e−βnε =

n=0

1 1 − e−βε

(9.27)

and the average energy at T of the photons in this mode is given by nε = −

ε 1 ∂z(ε) = βε z(ε) ∂β e −1

For the occupation factor of this mode, one ﬁnds again the Bose-Einstein distribution in which µ = 0, i.e. f (ε) =

1 −1

(9.28)

eβε

[Note that in the study of the speciﬁc heat of solids (see § 2.4.5) a factor similar to (9.28) appeared : in the Einstein and Debye models the quantized solid vibrations are described by oscillators. Changing the vibration

harmonic

state of the solid from n + 12 ω to n + 12 + 1 ω is equivalent to creating a quasi-particle, called a phonon, which follows the Bose-Einstein statistics. The number of phonons is not conserved]. If now one accounts for all the photons modes, for a macroscopic cavity volume p), D(ε) such that the Ω, one can deﬁne the densities of states Dk (k), Dp ( number dn of modes with a wave vector between k and k + dk is equal to dn : Ω 3 d k = Dk (k)d3k (2π)3 Ω p = Dp ( p)d3 p dn = 2 · 3 d3 h dn = 2 ·

that is, Dk (k) =

Ω 4π 3

,

Dp ( p) =

2Ω h3

(9.29)

The factor 2 expresses the two possible values of the spin of a photon of ﬁxed k (or p). Since the energy ε only depends of the modulus p, one obtains D(ε) from Ω Ω 4πp2 dp = 8π 3 3 ε2 dε 3 h h c 8πΩ 2 D(ε) = 3 3 ε h c

D(ε)dε = 2

(9.30)

Bose-Einstein Distribution of Photons

209

In this “large volume limit”, one can express the thermodynamical parameters. The thermodynamical potential to be considered here is the free energy F (since µ = 0) ∞ ln(1 − e−βε )D(ε)dε (9.31) F = −kB T ln Z = kB T 0

The internal energy is written ∞ U= εD(ε)f (ε)dε 0 ∞ ∞ 8πΩ ε3 8πΩ hν 3 dν = dε = h3 c3 eβε − 1 c3 eβhν − 1 0 0 i.e.,

(9.32)

∞

Ωu(ν)dν

U= 0

with u(ν) =

8πh ν3 c3 eβhν − 1

(9.33)

The photon spectral density in energy u(ν) is deﬁned from the energy contribution dU of the photons present inside the volume Ω, with a frequency between ν and ν + dν : dU = Ωu(ν)dν

(9.34)

The expression (9.33) of u(ν) is called the Planck law . The total energy for the whole spectrum is given by ∞ 3 8πΩ x dx U = 3 3 (kB T )4 h c ex − 1 0

(9.35)

where the dimensionless integral is equal to Γ(4)ζ(4) with Γ(4) = 3! and π4 ζ(4) = (see the section “Some useful formulae” in this book). Introducing 90 = h/2π, the total energy becomes U=

π 2 (kB T )4 Ω 15 (c)3

(9.36)

The free energy U and the internal energy F are easily related : in the integration by parts of (9.31), the derivation of the logarithm introduces

210

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

the Bose-Einstein distribution, the integration of D(ε) provides a term in ε3 /3 = ε . ε2 /3. This exactly gives F =−

π 2 (kB T )4 U =− Ω 3 45 (c)3

[to be compared to (9.11), which is valid for material particles with a density of states in ε1/2 , whereas the photons density varies in ε2 ]. Since dF = −SdT − P dΩ

(9.37)

one deduces Å S=− Å P =− P =

∂F ∂T ∂F ∂Ω

ã = Ω

ã

Å ã 4π 2 kB T 3 kB Ω 45 c

=− T

F Ω

π 2 (kB T )4 45 (c)3

(9.38) (9.39)

(9.40)

The pressure P created by the photons is called the radiation pressure.

9.2.3

Black Body Deﬁnition and Spectrum

The results obtained above are valid for a closed cavity, in which a measurement is strictly impossible because there is no access for a sensor. One is easily convinced that drilling a small hole into the cavity will not perturb the photons distribution but will permit measurements of the enclosed thermal radiation, through the observation of the radiation emitted by the hole. Besides any radiation coming from outside and passing through this hole will be trapped inside the cavity and will only get out after thermalization. This system is thus a perfect absorber, whence its name of black body. In addition we considered that thermal equilibrium at temperature T is reached between the photons and the cavity. In such a system, at steady-state, for each frequency interval dν, the energy absorbed by the walls exactly balances their emitted energy, which is found in the photon gas in this frequency range, that is, Ωu(ν)dν. Therefore the parameters we obtained previously for the photon gas are also those characterizing the thermal emission from matter at temperature T , whether this emission takes place in a cavity or not, whether this matter is in contact with vacuum or not. We will then be able to compare the laws already stated, and in the ﬁrst place the Planck law, to experiments performed on thermal radiation.

Bose-Einstein Distribution of Photons

211

The expression (9.33) of u(ν) provides the spectral distribution of the thermal radiation versus the parameter β, i.e. versus the temperature T . For any T , u(ν) vanishes for ν = 0 and tends to zero when ν tends to inﬁnity. The variable is in fact βhν and one ﬁnds that the maximum of u(ν) is reached for βhνmax = 2.82, that is, νmax = 2.82

kB T h

(9.41)

Relation (9.41) constitutes the Wien law. When T increases, the frequency of the maximum increases and for T1 < T2 the whole curve u(ν) for T2 is above that for T1 (Fig. 9.1). u(ν)

J.s.m−3

5.10−15 T1 < T2 < T3

T2

T3

=

60

00

K

T1 0

ν 10 Hz 14

4

Fig. 9.1: Spectral density in energy of the thermal radiation for several temperatures.

For small values of ν, u(ν) kB T · for ν large, u(ν)

8π 2 ν c3

8πh 3 −βhν ν e c3

(9.42) (9.43)

At the end of the 19th century, photometric measures allowed one to obtain the u(ν) curves with a high accuracy. Using the then available MaxwellBoltzmann statistics, Lord Rayleigh and James Jeans had predicted that the average at temperature T of the energy of oscillators should take the form (9.42), in kB T ν 2 . If this latter law does describe the low-frequency behavior of the thermal radiation, it predicted an “ultraviolet catastrophe” [u(ν) would be an increasing function of ν, thus the U V and X emissions should be huge]. Besides, theoretical laws had been obtained by Wilhem Wien and experimentally conﬁrmed by Friedrich Paschen : an exponential behavior at

212

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

high frequency like in (9.43) and the law (9.41) of the displacement of the distribution maximum. It was the contribution of Max Planck in 1900 to guess the expression (9.33) of u(ν), assuming that the exchanges of energy between matter and radiation can only occur through discrete quantities, the quanta. Note that the ideas only slowly clariﬁed until the advent in the mid 1920’s of Quantum Mechanics, such as we know it now. It it remarkable that the paper by Bose, proposing the now called “Bose-Einstein statistics,” preceded by a few months the formulation by Erwin Schroedinger of its equation ! The universe is immersed within an infrared radiation, studied in cosmology, the distribution of which very accurately follows a Planck law for T = 2.7 K. It is in fact a “fossil” radiation resulting from the cooling, by adiabatic expansion of the universe, of a thermal radiation at a temperature of 3000 K which was prevailing billions of years ago : according to expression (9.38) of the entropy, an adiabatic process maintains the product ΩT 3 . The universe radius has thus increased of a factor 1000 during this period. The sun is emitting toward us radiation with a maximum in the very near infrared, corresponding to T close to 6000 K. A “halogen” bulb lamp is emitting a radiation corresponding to the temperature of its tungsten ﬁlament (around 3000 K), which is immersed in a halogen gas to prevent its evaporation. We ourselves are emitting radiation with a maximum in the far infrared, corresponding to our temperature close to 300 K.

9.2.4

Microscopic Interpretation of the Bose-Einstein Distribution of Photons

Another schematic way to obtain the Bose-Einstein distribution, and consequently, the Planck law, consists in considering that the walls of the container are made of atoms with only two possible energy levels ε0 and ε1 separated by ε1 − ε0 = hν, and in assuming that the combined system (atoms-photons of energy hν) is at thermal equilibrium (Fig. 9.2). It is possible to rigorously calculate the transition probabilities between ε0 and ε1 (absorption), or between ε1 and ε0 (emission) under the eﬀect of photons of energy hν. This is done in advanced Quantum Mechanics courses. Here we will limit ourselves to a more qualitative argument, admitting that in Quantum Mechanics a transition is expressed by a matrix element, which is the product of a term 1|V |0 speciﬁc of the atoms with another term due to the photons. Besides, we will admit that an assembly of n photons of energy hν is described by a harmonic oscillator, with energy levels splitting equal to hν, lying in the state |n. The initial state contains N0 atoms in the fundamental state, N1 atoms in the state ε1 , and n photons of energy hν.

Bose-Einstein Distribution of Photons

213

BEFORE

AFTER

N1 hν

E1

N0

E0

N1

E1

N0

E0

N1

E1

N1 + 1 absorption N0 − 1

N1 − 1 hν spontaneous emission N0 + 1

N1 − 1 hν

hν N0

E0

induced emission

N0 + 1

Fig. 9.2: Absorption, spontaneous and induced emission processes for an assembly of two-level atoms in thermal equilibrium with photons.

In the absorption process (Fig. 9.2 top), the number of atoms in the state of energy ε0 becomes N0 − 1, while N1 changes to N1 + 1 ; a photon is utilized, thus the photons assembly shifts from |n to |n − 1. This is equivalent to considering the action on the state |n of the photon annihilation operator, i.e., according to the Quantum Mechanics course, a|n =

√ n |n − 1

(9.44)

The absorption probability per unit time P is proportional to the number of atoms N0 capable of absorbing a photon, to the square of the matrix element appearing in the transition, and is equal, to a multiplicative constant, to : Pa = N0 |1|V |0|2 |n − 1|a|n|2 = N0 |1|V |0| n 2

(9.45) (9.46)

This probability is, as expected, proportional to the number n of present photons.

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Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

In the emission process (Figs. 9.2 middle and bottom), using a similar argument, since the number of photons is changed from n to n + 1, one considers the action on the state |n of the photon creation operator : √ (9.47) a+ |n = n + 1 |n + 1 The emission probability per unit time Pe is proportional to the number of atoms N1 that can emit a photon, to the square of the matrix element of this transition and is given by Pe = N1 |0|V |1|2 |n + 1|a+ |n|2 = N1 |0|V |1| (n + 1) 2

(9.48) (9.49)

We ﬁnd here that Pe is not zero even when n = 0 ( spontaneous emission), but that the presence of photons increases Pe (induced or stimulated emission) : another way to express this property is to say that the presence of photons “stimulates” the emission of other photons. The induced emission was introduced by A. Einstein in 1916 and is at the origin of the laser eﬀect. Remember that the name “laser” stands for Light Ampliﬁcation by Stimulated Emission of Radiation. In equilibrium at the temperature T , n photons of energy hν are present : the absorption and emission processes exactly compensate, thus N0 n = N1 (n + 1)

(9.50)

Besides, for the atoms in equilibrium at T , the Maxwell-Boltzmann statistics allows one to write N1 = e−βhν (9.51) N0 One then deduces the value of n n=

1 eβhν

−1

(9.52)

which is indeed the Bose-Einstein distribution of photons. This yields u(ν, T ) through the same calculation of the photons modes as above [formulae (9.29) to (9.33)]. This approach, which is based on advanced Quantum Mechanical results, has the advantage of introducing the stimulated emission concept and the description of an equilibrium from a balance between two phenomena.

9.2.5

Photometric Measurements : Deﬁnitions

We saw that a “black body” is emitting radiation corresponding to the thermal equilibrium in the cavity at the temperature T . A few simple geometrical

Bose-Einstein Distribution of Photons

215

considerations allow us to relate the total power emitted by a black body to its temperature and to introduce the luminance, which expresses the “color” of a heated body : we will then be able to justify that all black bodies at the same given temperature T are characterized by the same physical parameter and thus that one can speak of THE black body. Let us ﬁrst calculate the total power P emitted through the black-body aperture in the half-space outside the cavity. We consider the photons of frequency ν, to dν, which travel through the aperture of surface dS and have a velocity directed toward the angle θ to dω with the normal n to dS. The number of these photons which pass the hole during dt are included in a cylinder of basis dS and height c cos θdt (Fig. 9.3). They carry an energy

dω θ dS

n

cdt

Fig. 9.3: Radiation emitted through an aperture of surface dS in a cone dω around the direction θ.

dω c cos θ dt dS u(ν)dν (9.53) 4π The total power P radiated into the half-space, by unit time, through a hole of unit surface is given by 2π π2 dϕ ∞ cos θ sin θdθ c u(ν, T )dν (9.54) P = 4π 0 0 0 d2 P · dt =

P =

π 2 (kB T )4 cU = = σT 4 4Ω 60 3 c2

(9.55)

This power varies like T 4 , this is the Stefan-Boltzmann law. The Stefan constant σ is equal to 5.67 × 10−8 watts · m−2 · kelvins−4 . At 6000 K, the approximate temperature of the sun surface, one gets P = 7.3 × 107 W· m−2 ; at 300 K one ﬁnds 456 W·m−2 . Another way to handle the question is to analyze the collection process of the radiation issued from an arbitrary source by a detector. It is easy to

216

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

understand that the collected signal ﬁrst depends on the spectral response of the detector : for instance, a very speciﬁc detector is our eye, which is only sensitive between 0.4 µm and 0.70 µm, with a maximum of sensitivity in the yellow-green at 0.55 µm. Therefore speciﬁc units have been deﬁned in visual photometry, which take into account the physiological spectral response. But the signal also depends on geometrical parameters, such as the detector surface dS , the solid angle under which the detector sees the source or, equivalently, the source surface dS and the source-detector distance together with their respective directions (Fig. 9.4). A physical parameter, the luminance L, characterizes the source : in the visible detection range the colors of two sources of diﬀerent luminances are seen as diﬀerent (a red-hot iron does not have the same aspect as the sun or as an incandescent lamp ; a halogen lamp looks whiter that an electric torch). To be more quantitative, in energetic photometry one deﬁnes the inﬁnitesimal power d2 P received by the elementary surface dS of a detector from the surface dS of a source (Fig. 9.4) by the equation d2 P = LdS cos θdωdν = LdS

cos θ cos θ dS dν = LdS cos θ dω dν r2

dS

dS dω M

(9.56)

θ n

dω θ

M

n

Fig. 9.4: Deﬁnition of the geometrical parameters of the source and detector. The surface element of the source is centered on M , the detection element on M , with M M = r ; θ is the angle between M M and the normal n to dS, dω the solid angle under which the detector is seen from the source dS. One notices that the deﬁnition is fully symmetrical and that alternatively one can introduce dS , the solid angle dω under which the source is seen from the detector, the angle θ between M M and the normal to dS in M . The luminance depends of T , ν and possibly of θ, the direction of the emission, that is, L(ν, T, θ). When the source is a black body of surface dS in thermal equilibrium at the temperature T , the power emitted toward dS in the frequency range (ν, ν+dν)

Bose-Einstein Distribution of Photons

217

is calculated from (9.53) d2 P =

c u(ν, T ) dS cos θ dω dν 4π

(9.57)

The comparison between (9.56) and (9.57) allows one to express the black body luminance L0 versus u(ν, T ) : c u(ν, T ) (9.58) L0 (ν, T ) = 4π Thus all black bodies at the same temperature T have the same luminance, which allows one to deﬁne the black body. It is both the ideal source and detector since, once the equilibrium has been reached, it absorbs any radiation that it receives and re-emits it. Its luminance does not depend on the direction θ of emission. Formulae (9.56) and (9.57) also express that the surface which comes into play is the apparent surface, i.e., the projection dS cos θ of the emitting surface on the plane perpendicular to M M . For this whole surface c u(ν, T ) : this is the Lambert law the luminance is uniform and equal to 4π which explains why the sun, a sphere radiating like the black body, appears to us like a planar disk. The heated bodies around us are generally not strictly black bodies, in the meaning that (9.58) is not fulﬁlled : their luminance(L(ν, T, θ)) not only depends on the frequency and the temperature, but also on the emission direction. Besides, we saw that the black body is “trapping” any radiation which enters the cavity : it thus behaves like a perfect absorber, which is not the most general case for an arbitrary body.

9.2.6

Radiative Balances

Consider an arbitrary body, exchanging energy only through radiation and having reached thermal equilibrium : one can only express a balance stating that the absorbed power is equal to the emitted power (in particular a transparent body neither absorbs nor emits energy). Assume that the studied body is illuminated by a black body and has an absorption coeﬃcient a(ν, T, θ) (equal to unity for the black body). If the studied body does not diﬀuse light, that is, radiation is only reemitted in the image direction of the source, and if it remains at the frequency it was absorbed, one obtains the relation c u(ν, T )a(ν, T, θ) = L(ν, T, −θ) (9.59) L0 (ν, T )a(ν, T, θ) = 4π emission absorption after simplifying the geometrical factors. Relation (9.59) links the black body luminance L0 (ν, T ) to that of the considered body, of absorption coeﬃcient a(ν, T, θ). The black body luminance is

218

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

larger than, or equal to, that of any thermal emitter at the same temperature, the black body is the best thermal emitter. This is the Kirchhoﬀ law. In the case of a diﬀusing body (reemitting radiation in a direction other than −θ) or of a ﬂuorescent body (reemitting at a frequency diﬀerent from that of its excitation), one can no longer express a detailed balance, angle by angle or frequency by frequency, between the ambient radiation and the body radiation. However, in steady-state regime, the total absorbed power is equal to the emitted power and the body temperature remains constant. The applicable relation is deduced from (9.59) after integration over the angles and/or over the frequencies. One can also say that the power, assumed to be emitted by a black body, received by the studied body at thermal equilibrium at T , is either absorbed (and thus reemitted since T remains constant), or reﬂected. For instance for a diﬀusing body one writes : received power = emitted power + reﬂected power which implies for the luminances : L0 (ν, T ) = L(ν, T ) + [L0 (ν, T ) − L(ν, T )] L0 (ν, T ) = L(ν, T ) + [1 − a(ν, T )]L0 (ν, T )

(9.60)

This is another way to obtain (9.59). These types of balances are comparisons between the body luminance L and the black body luminance L0 : if the interpretation of the L0 expression requires Quantum Mechanics, the comparisons of luminances are only based on balances and were discovered during the 19th century. More generally speaking, the exchanges of energy are taking place according to three processes : conduction, convection and radiation. Whereas the ﬁrst two mechanisms require a material support, radiation which occurs through photons can propagate through vacuum : it is thanks to radiation that we are in contact with the universe. Through the intersideral vacuum, stars are illuminating us and we can study them by analyzing the received radiation. Our star, the sun, is sending a power of the order of a kW to each 1 m2 surface of earth. The researches on solar energy try to make the best use of this radiation, using either thermal sensors, or photovoltaic cells (“solar cells”) (see §8.5.2) which directly convert it into electricity.

9.2.7

Greenhouse Eﬀect

Before interpreting the radiative balance of the earth, let us consider the greenhouse eﬀect which allows one to maintain inside a building, made of

Bose-Einstein Distribution of Photons

219

glass formerly, often of plastics nowadays, a temperature higher than the one outside and thus favors particular vegetables or ﬂowers cultivation (Fig. 9.5). The origin of the eﬀect arises from the diﬀerence in absorption properties of the building walls for the visible and infrared radiations. In the visible range, where the largest part of the radiation emitted by the sun is found, the walls are transparent, so that the sun radiation can reach the ground, where it is absorbed. The ground in turn radiates like a black body, with its frequency maximum in the infrared [Wien law, equation (9.41)]. Now the walls are absorbing in the infrared, they reemit a part of this radiation toward the ground, the temperature of which rises, and so on. At equilibrium the ground temperature is thus TG. If a steady state has been reached, the power getting into the greenhouse, i.e., P0 , is equal to the one that is getting out of it :

P0

Wall in glass or plastics a

TG

Ground TG

Fig. 9.5: Principle of the greenhouse eﬀect.

P0 = σSTG4 (1 − a) if S is the ground surface and a the absorption coeﬃcient, by the walls, of the ground infrared radiation. In the absence of greenhouse, the ground temperature would be TG , such that P0 = σSTG4 The temperature TG inside the greenhouse is higher than TG since TG =

TG (1 − a)1/4

Here we assumed a perfect transparency of the walls in the visible and no reﬂection by the walls both in the visible and the infrared. “The greenhouse eﬀect” related to a possible climate global warming is an example of radiative balance applied to the earth. The earth is surrounded by a layer of gases, some of which corresponding to molecules with a permanent dipolar moment, due to their unsymetrical arrangement of atoms : this is the case for H2 O, CO2 , O3 , etc. Thus the vibration and rotation spectra of these molecules evidence an infrared absorption. The sun illuminates the earth with thermal radiation, the maximum of its spectrum being located in the visible [Eq. (9.41)] ; this radiation is warming

220

Chapter 9. Bosons : Helium 4, Photons, Thermal Radiation

the ground. The earth, of temperature around 300 K, reemits thermal radiation toward space, the frequency maximum of this reemission is in the infrared. A fraction of the earth radiation is absorbed by the gases around the earth, which in turn reemit a fraction of it, and so on. Finally, according to the same phenomenon as the one described above for a greenhouse, the earth receives a power per unit surface larger than the one it would have reached it if there had only been vacuum around it ; its average temperature is 288 K (instead of 255 K if it had been surrounded by vacuum). The global climate change, possibly induced by human activity, is related to an increase in concentration of some greenhouse eﬀects gases (absorbing ﬂuorocarboned compounds, which are now forbidden for use, and mainly CO2 ). If the absorption coeﬃcient increases, the ground temperature will vary in the same way. This will produce eﬀects on the oceans level, etc. . . 1 More advanced studies on the greenhouse eﬀect account for the spectra of absorption or transparency of the gases layer around the earth according to their chemical species, etc. In our daily life, we are surrounded by radiating objects like heating radiators, incandescent lamps. On the contrary we wish other objects to maintain their energy without radiating : this is the case, in particular, for thermos bottles, Dewar cans. We foresee here a very vast domain with technological, ecological and political involvements !

1 The Intergovernmental Panel on Climate Change, established by the World Meteorological Organization and the United Nations Environment program, is assessing the problem of potential global climate change. An introduction brochure can be downloaded from its website (www.ipcc.ch).

Summary of Chapter 9 The chemical potential µ of material non-interacting particles following the Bose-Einstein statistics is located below their fundamental state ε1 . The occupation factor by material bosons of an energy level ε is equal to f (ε) =

1 eβ(ε−µ) − 1

At given density N/Ω, below a temperature TB , which is a function of N/Ω, the Bose condensation takes place : a macroscopic number of particles then occupies the fundamental state ε1 , that is, the same quantum state. The thermodynamical properties in this low temperature range are mainly due to the particles still in the excited states. The superﬂuidity of helium 4, the superconducting state of some solids, are examples of systems in the condensed Bose state, which is observed for T < TB . Photons are bosons in non-conserved number, their Bose-Einstein distribution is given by 1 f (ε) = βε e −1 The Planck law expresses the spectral density in energy or frequency of photons in an empty cavity of volume Ω, in the frequencies range [ν, ν + dν] : dU = Ωu(ν)dν = Ω

8πh ν 3 dν c3 eβhν − 1

This law gives the spectral distribution of the “black body” radiation, i.e., of thermal radiation at temperature T = 1/kB β. The black body is the perfect thermal emitter, its luminance is proportional to u(ν). The thermal emission of a body heated at temperature T corresponds to a luminance at maximum equal to that of the black body. 221

222

Summary of Chapter 9

The greenhouse eﬀect arises from the properties of the walls enclosing the greenhouse : higher transparency in the visible and larger infrared absorption. This phenomenon is also applicable to the radiation balance of the earth.

General Method for Solving Exercises and Problems At the risk of stating obvious facts, we recall here that the text of a statistical physics exercise or problem is very precise, contains all the required information and that each word really matters. Therefore it is recommended to ﬁrst read the text with the highest attention. The logic of the present course will always be apparent, i.e, you have to ask yourself the two following questions, in the order given here : 1. What are the microstates of the studied system ? 2. Which among these microstates are achieved in the problem conditions ? 1. The ﬁrst question is generally solved through Quantum Mechanics, which reduces to Classical Mechanics in special cases (the ideal gas, for example). Mostly this is equivalent to the question : “What are the energy states of the system ?” In fact they are often speciﬁed in the text. At this stage there are three cases : – distinguishable particles (for example of ﬁxed coordinates) : see the ﬁrst part of the course – classical mobile particles : see the chapter on the ideal gas – indistinguishable particles : see the second part of the course, in which the study is limited to the case of independent particles.

223

224

General Method for Solving Exercises and Problems

2. The second question speciﬁcally refers to Statistical Physics. According to the physical conditions of the problem one will work in the adapted statistical ensemble : – isolated system, ﬁxed energy and ﬁxed particles number : microcanonical ensemble – temperature given by an energy reservoir and ﬁxed number of particles : canonical ensemble – system in contact with an energy reservoir (which dictates its temperature) and a particles reservoir (which dictates its chemical potential) : grand canonical ensemble The systems of indistinguishable particles obeying the Pauli principle and studied in the framework of this course are solved in the grand-canonical ensemble. Their chemical potential is then set in such a way that the average number of particles coincides with the real number of particles given by the physics of the system.

Units and physical constants (From R. Balian, “From microphysics to macrophysics : methods and applications of statistical physics”, Springer Verlag Berlin (1991)) We use the international system (SI), legal in most countries and adopted by most international institutions. Its fundamental units are the meter (m), the kilogram (kg), the second (sec), the ampere (A), the kelvin (K), the mole (mol), and the candela (cd), to which we add the radian (rad) and the steradian (sr). The derived units called by speciﬁc names are : the hertz (Hz = s−1 ), the newton (N = m kg sec−2 ), the pascal (Pa = N m−2 ), the joule (J = N m), the watt (W = J sec−1 ), the coulomb (C = A s), the volt (V = W A−1 ), the farad (F = C V−1 ), the ohm (Ω = V A−1 ), the siemens (S = A V−1 ), the weber (Wb = V s), the tesla (T = Wb m−2 ), the henry (H = Wb A−1 ), the Celsius degree (C), the lumen (lm = cd sr), the lux (lx = lm m−2 ), the becquerel (Bq = sec−1 ), the gray (Gy = J kg−1 ) and the sievert (Sv = J kg−1 ). The multiples and submultiples are indicated by the preﬁxes deca (da = 10), hecto (h = 102 ), kilo (k = 103 ) ; mega (M = 106 ), giga (G = 109 ), tera (T = 1012), peta (P = 1015 ), exa (E = 1018 ) ; deci (d = 10−1 ), centi (c = 10−2 ), milli (m = 10−3 ) ; micro (µ = 10−6 ), nano (n = 10−9 ), pico (p = 10−12 ), femto (f = 10−15 ), atto (a = 10−18 ).

225

226

Units and physical constants

Constants of electromagnetic units light velocity Planck constant Dirac constant Avogadro number Atomic mass unit neutron and proton masses electron mass Elementary charge Faraday constant Bohr magneton nuclear magneton Fine structure constant Hydrogen atom : Bohr radius binding energy Rydberg constant Boltzmann constant molar constant of the gas Normal conditions : pressure temperature

molar volume Gravitation constant gravity acceleration Stefan constant

µ0 = 4π × 10−7 N A−2 (deﬁnes the ampere) ε0 = 1/µ0 c2 , 1/4πε0 9 × 109 N m2 C−2 c = 299 792 458 m sec−1 (deﬁnes the meter) c 3 × 108 m sec−1 h = 6.626 × 10−34 J sec = h/2π 1.055 × 10−34 J sec N 6.022 × 1023 mol−1 (by deﬁnition, the mass of one mole of 12 C is 12 g) 1 u=1 g/N 1.66 × 10−27 kg (or dalton or amu) mn 1.0014mp 1.0088 u m 1 u/1823 9.11 × 10−31 kg e 1.602 × 10−19 C N e 96 485 C mol−1 µB = e/2m 9.27 × 10−24 J T−1 e/2mp 5 × 10−27 J T−1 e2 1 α= 4πε0 c 137 4πε0 2 = 0.53 Å mcα me2Å ã2 2 m e2 E0 = = 2 2ma20 2 4πε0 13.6 eV R∞ = E0 /hc 109 737 cm−1 kB 1.381 × 10−23 J K−1 R = N kB 8.316 J K−1 mol−1 a0 =

1 atm = 760 torr = 1.01325× 105 Pa Triple point of water : 273.16 K ( deﬁnition of the kelvin) or 0.01C (deﬁnition of the Celsius scale) 22.4 × 10−3 m3 mol−1 G 6.67 × 10−11 m3 kg−1 sec−2 g 9.81 m sec−2 π2 k4 603 c2 5.67 × 10−8 W m−2 K−4 σ=

Units and physical constants

227

Deﬁnition of the photometric units

A light power of 1 W, at the frequency of 540 THz, is equivalent to 683 lm

Energy units and equivalences :

1 erg = 10−7 J (not IS) 1 kWh=3.6×106 J 1 eV ↔ 1.602 × 10−19 J ↔ 11 600 K

electrical potential (electron-volt) heat (calorie) chemical binding temperature (kB T ) mass (mc2 ) wave number (hc/λ) frequency (hν)

1 cal = 4.184 J (not SI ; speciﬁc heat of 1 g water) 23 kcal mol−1 ↔ 1 eV (not SI) 1 290 K ↔ 40 eV (standard temperature) 9.11 × 10−31 kg ↔ 0.511 MeV (electron at rest) 109 700 cm−1 ↔ 13.6 eV (Rydberg) 3.3 × 1015 Hz ↔ 13.6 eV

It is useful to remember these equivalencies to quickly estimate orders of magnitude. Various non-SI units

1 ångström (Å)=10−10 m (atomic scale) 1 fermi (fm)=10−15 m (nuclear scale) 1 barn (b)=10−28 m2 1 bar=105 Pa 1 gauss (G)=10−4 T 1 marine mile=1852 m 1 knot=1 marine mile per hour 0.51 m sec−1 1 astronomical unit (AU) 1.5 × 1011 m (Earth-Sun distance) 1 parsec (pc) 3.1 × 1016 m (I UA/sec arc) 1 light-year 0.95 × 1016 m

Data on the Sun

Radius 7 × 108 m=109 earth radii Mass 2 × 1030 kg Average density 1.4 g cm−3 Luminosity 3.8 × 1026 W

A few useful formulae Normalization of a Gaussian function : … +∞ π −ax2 dx e = a −∞ The derivation of this formula with respect to a yields the moments of the Gauss distribution.

Euler gamma function :

∞

Γ(t) ≡

xt−1 e−x dx = (t − 1)Γ(t − 1)

0

∞

(1 − x)s−1 xt−1 dx =

0

Γ(t)Γ(1 − t) = Stirling formula :

π sin πt

,

Γ(s)Γ(t) Γ(s + t)

Γ

Å ã √ 1 = π 2

√ t! = Γ(t + 1) ∼ tt e−t 2πt t→∞

Binomial series : (1 + x)t =

∞ n x

∞ (−x)n Γ(n − t) Γ(t + 1) = n! Γ(t + 1) − n) n=0 n! Γ(−t) n=0

,

Poisson formula : +∞

f (n) =

n=−∞

+∞

f˜(2πl) ≡

l=−∞

+∞ l=−∞

229

+∞

−∞

dxf (x)e2πilx

|x| < 1

230

A few useful formulae

Euler-Maclaurin formula : ε 1 ε3 1 a+ε f (x) |a+ε dxf (x) ≈ [f (a) + f (a + ε)] − f (x) |a+ε + +... a a ε a 2 12 720 ε 7ε3 ≈ f (a + 12 ε) + f (x) |a+ε f (x) |a+ε − +... a a 24 5760 This formula allows one to calculate the diﬀerence between an integral and a sum over n, for a = nε. Constants e 2.718, π 3.1416 1 γ ≡ lim(1 + ... + − ln n) 0.577 Euler constant. n Riemann zeta function : ∞ t−1 ∞ 1 x dx = Γ(t)ζ(t) ζ(t) ≡ , t n ex − 1 0 n=1

∞

0

t ζ

1.5 2.612

2

xt−1 dx = (1 − 2−t+1 )Γ(t)ζ(t) ex + 1

2.5 1.341

1 2 6π

Dirac distribution : 1 2π

+∞

−∞

3 1.202

3.5 1.127

4 1 4 90 π

5 1.037

dx eixy/a = δ(y/a) = |a|δ(y)

+∞ 1 2πi y/a e = δ(y/a) |a|

=−∞

lim

t→∞

sin tx 1 − cos tx = lim = πδ(x) t→∞ x tx2

f (x)δ(x) = f (0)δ(x)

,

f (x)δ (x) = −f (0)δ(x) + f (0)δ (x)

If f (x) = 0 at the x = xi points, one has 1 δ(x − xi ) δ[f (x)] = |f (xi )| i

Exercises and Problems A course can only be fully understood after practice through exercises and problems. There are many excellent collections of classical exercises in Statistical Physics. In what follows, we reproduce original texts which were recently given as examinations at Ecole Polytechnique. An examination of Statistical Physics at Ecole Polytechnique usually consists in an exercise and a problem. As examples, you will ﬁnd here the exercises given in 2000, 2001 and 2002 and the problems of 2001 and 2002 which were to be solved by the students who had attended the course presented in this book. Electrostatic Screening : exercise 2000 Magnetic Susceptibility of a “Quasi-1D” Conductor : exercise 2001 Entropies of the HC Molecule : exercise 2002 Quantum Boxes and Optoelectronics : problem 2001 Physical Foundations of Spintronics : problem 2002

231

Exercise 2000 : Electrostatic Screening

233

Exercise 2000 : Electrostatic Screening I. Let us consider N free fermions of mass me and spin 1/2, without any mutual interaction, contained in a macroscopic volume Ω maintained at temperature T . I.1. Recall the equation relating N/Ω to the chemical potential µ of these fermions through the density of states in k (do not calculate the integral). I.2. In the high temperature and low density limit recall the approximate integral expression relating N/Ω to µ (now calculate the integral). I.3. The same fermions, at high temperature and small density, are now submitted to a potential energy V (r). Express the total energy of one such particle versus k. Give the approximate expression relating N to µ and V (r). I.4. Deduce that the volume density of fermions is given by n(r) = n0 exp(−βV (r))

,

where β = 1/kB T and n0 is a density that will be expressed as an integral. I.5. Show that at very small potential energy, such that βV (r) 1 for any r, then n0 = N/Ω. II. From now on, we consider an illuminated semiconductor, of volume Ω, maintained at temperature T . It contains N electrons, of charge −e, in its conduction band together with P = N holes (missing electrons) in its valence band. The holes will be taken as fermions, of spin 1/2, mass mh and charge +e. We assume that the high temperature limit is realized. Assume that a ﬁxed point charge Q is present at the coordinates origin, which, if alone, would create an electrostatic potential : Φext (r) =

Q 4πε0 εr r

where εr is the relative dielectric constant of the considered medium. In fact, the external charge Q induces inhomogeneous charge densities −en(r) and +ep(r) inside both fermion gases. These induced charges in turn generate electrostatic potentials, which modify the densities of mobile charges, and so on. You are going to solve this problem by a self-consistent method in the following way : let Φtot (r) be the total electrostatic potential, acting on the particles.

234

Exercises and Problems

II.1. Express the potential energies, associated with Φtot (r), to which the electrons and the holes are respectively subjected. Deduce from I. the expressions n(r) and p(r) of the corresponding densities of mobile particles. II.2. The electrostatic potential Φtot (r) is a solution of the Poisson equation, in which the charge density includes all the charges of the problem. Admit that the charge density Qδ(r) is associated with the point charge Q located at r, where δ(r) is the Dirac delta distribution. Show that, in the high temperature limit and to the lowest order in Φtot (r), the total electrostatic potential is solution of ∆Φtot (r) =

Φtot (r) Qδ(r) − λ2 ε0 εr

where λ is a constant that will be expressed versus N/Ω and T . What is the unit of λ ? II.3. Estimate λ for N/Ω = 1022 m−3 , T = 300 K, εr = 12.4. II.4. The solution of this equation is Φtot (r) =

Q exp(−r/λ) 4πε0 εr r

Indicate the physical implications of this result in a few sentences.

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” I. Consider an assembly of N magnetic ions at temperature T inside a volume Ω. These ions are independent and distinguishable. The electronic structure of an ion consists in a fundamental non-degenerate level |0, of energy ε0 , and a doubly degenerate excited state of energy ε0 + λ. Under the eﬀect of a static and uniform magnetic ﬁeld B, the fundamental level remains at the energy ε0 while the excited level splits into two sublevels |+ and |− of respective energies ε0 + λ + γB and ε0 + λ − γB. I.1. Without any calculation, plot the location of the states |0, |+ and |− versus the applied magnetic ﬁeld, in the following special cases : a) λ = 0 , kB T γB b) λ kB T γB c) kB T λ γB d) γB λ kB T

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” Conductor

235

In each case indicate which states are occupied at temperature T . I.2. Calculate the system free energy, assuming that the only degrees of freedom are those, of electronic origin, related to the three levels |0, |+ and |−. ∂F and Deduce the algebraic expressions of the system magnetization M = − ∂B M of its susceptibility χ = limB→0 (Ω is the volume). ΩB I.3. On the expressions of the magnetization and of the susceptibility obtained γB = in I.2. only discuss the limit cases a), b), c), d) of question I.1 (take kB T x). Verify that the obtained results are in agreement with the qualitative considerations of I.1.

Susceptibility

I.4. The ﬁgure below represents the variation versus temperature of the paramagnetic susceptibility of the organic “quasi-one-dimensional” compound [HN(C2 H5 )3 ](TCNQ)2 . Using the results of the present exercise can you qualitatively interpret the shape of this curve ?

Temperature (K)

Fig. 1: Magnetic susceptibility of the quasi-one-dimensional organic compound versus temperature. II. Let a paramagnetic ion have a total spin S. You will admit that, in presence the levels of this ion take the energies of a static and uniform magnetic ﬁeld B, εl = lγB, where −S ≤ l ≤ S, with l = −S, (−S + 1), (−S + 2), . . . + S. Consider an ion at thermal equilibrium at the given temperature T . Calculate

236

Exercises and Problems

its free energy f and its average magnetic moment m = −

∂f . ∂B

Write the expressions of the magnetization in the following limit cases : high and low magnetic ﬁelds ; high and low temperatures. In each situation give a brief physical comment.

Exercise 2002 : Entropies of the HC Molecule 1. Let a three-dimensional ideal gas be made of N monoatomic molecules of mass m, conﬁned inside the volume Ω and at temperature T . Without demonstration recall the expressions of its entropy S3D and of its chemical potential µ3D . 2. For a two-dimensional ideal gas of N molecules, mobile on a surface A, of the same chemical species as in 1., give the expressions, versus A, m and T , of the partition function Z2D , the entropy S2D and the chemical potential µ2D . To obtain these physical parameters, follow the same approach as in the course. Which of these parameters determines whether a molecule from the ideal gas of question 1. will spontaneously adsorb onto surface A ? 3. Let us now assume that each of the N molecules from the gas is adsorbed on surface A. What is, with respect to the gas phase, the variation per mole of translation entropy if the molecules now constitute a mobile adsorbed ﬁlm on A ? 4. Numerically calculate the value of S3D (translation entropy) for one mole of HC, of molar mass M = 36.5, at 300K under the atmospheric pressure. Calculate S2D (translation entropy on a plane) for one mole of such a gas, adsorbed on a porous silicon surface : assume that, due to the surface corrugation, the average distance between molecules on the surface is equal to 1/30 of the distance in the gas phase considered above. 5. In fact, in addition to its translation degrees of freedom, the HC molecule possesses rotation degrees of freedom. Write the general expression of the partition function of N molecules of this gas versus ztr and zrot , the translation and rotation partition functions for a single molecule, accounting for the fact that the molecules are indistinguishable. The expressions of ztr and zrot will be calculated in the next questions. 6. In this question one considers a molecule constrained to rotate in a plane around a ﬁxed axis. Its inertia moment with respect to this axis is IA . The

Problem 2001 : Quantum Boxes and Optoelectronics

237

rotation angle φ varies from 0 to 2π, the classical momentum associated with φ is pφ . The hamiltonian for one molecule corresponding to this motion is given by p2φ h= 2IA The diﬀerential elementary volume of the corresponding phase space is dφ dpφ /h, where h is the Planck constant. Calculate the partition function z1rot . 7. In the case of the linear HC molecule, of inertia moment I, free to rotate in space, the rotation hamiltonian for a single molecule is written, in spherical coordinates, p2φ p2 h= θ + 2I 2I sin2 θ Here θ is the polar angle (0 ≤ θ ≤ π), φ the azimuthal angle (0 ≤ φ ≤ 2π) ; pθ (respectively pφ ) is the momentum associated with θ (respectively φ). The volume element of the phase space is now equal to dφ dpφ dθ dpθ /h2 . Calculate the corresponding partition function z2rot . Show that the dependence in temperature and in inertia moment obtained for z2rot is the expected one, when this expression of the rotation energy is taken into account. Verify that the expression found for z2rot is in agreement with the one given in the course for a linear molecule in the classical approximation (be careful, the HC molecule is not symmetrical). 8. Numerically calculate the rotation entropy by mole at 300K of a HC mole. Take I = 2.7 × 10−40 g · cm 2 .

Problem 2001 : Quantum Boxes and Optoelectronics For more clarity, comments are written in italics and questions in standard characters. Semiconducting quantum boxes are much studied nowadays, because of their possible applications in optoelectronics. The aim of the present problem is to study the rate of spontaneous emission and the radiative eﬃciency in quantum boxes : these are two important physical parameters for the elaboration of emission devices (quantum–boxes lasers). We will particularly focus on the experimental dependence of these parameters.

238

Exercises and Problems

Fig. 1 : Electron microscope image of quantum boxes of InAs in GaAs.

Quantum boxes are made of three-dimensional semiconducting inclusions of nanometer size (see Fig.1). In the studied case, the inclusions consist of InAs and the surrounding material of GaAs. The InAs quantum boxes conﬁne the electrons to a nanometer scale in three dimensions and possess a discrete set of electronic valence and conduction states ; they are often called “artiﬁcial atoms.” In the following we will adopt a simpliﬁed description of their electronic structure, as presented below in Fig. 2.

Fig. 2: Simpliﬁed model of the electronic structure of an InAs quantum box and of the ﬁlling of its states at thermal equilibrium at T = 0. We consider that the energy levels in the conduction band, labeled by a positive integer n, are equally spaced and separated by ωc ; the energy of the fundamental level (n = 0) is noted Ec . The nth level is 2(n + 1) times degenerate, where the factor 2 accounts for the spin degeneracy and (n + 1) for the orbital degeneracy. Thus a conduction state is identiﬁed through the set of three in-

Problem 2001 : Quantum Boxes and Optoelectronics

239

1 teger quantum numbers |n, l, s, with n > 0 and 0 ≤ l ≤ n, s = ± . The ﬁrst 2 valence states are described in a similar way, with analogous notations (Ev , ωv , n , l , s ).

I. Quantum Box in Thermal Equilibrium At zero temperature and thermal equilibrium, the valence states are occupied by electrons, whereas the conduction states are empty. We are going to show in this section that the occupation factor of the conduction states of the quantum box in equilibrium remains small for temperatures up to 300K. Therefore, only in this part, the electronic structure of the quantum box will be roughly described by just considering the conduction and valence band levels lying nearest to the band gap. Fig.3 illustrates the assumed electronic conﬁguration of the quantum box at zero temperature.

Fig. 3 : Electronic conﬁguration at T = 0. I.1. Give the average number of electrons in each of these two (spindegenerate) levels versus β = 1/kB T , where T is the temperature, and the chemical potential µ . I.2. Find an implicit relation deﬁning µ versus β, by expressing the conservation of the total number of electrons in the thermal excitation process. I.3. Show that, in the framework of this model, µ is equal to (Ec + Ev )/2 independently of temperature. I.4. Estimate the average number of electrons in the conduction level at 300K and conclude. For the numerical result one will take Ec − Ev = 1 eV. In optoelectronic devices, like electroluminescent diodes or lasers, electrons are injected into the conduction band and holes into the valence band of the semiconducting active medium : the system is then out of equilibrium. The return toward equilibrium takes place through electron transitions from the conduction band toward the valence band, which may be associated with light emission.

240

Exercises and Problems

Although the box is not in equilibrium, you will use results obtained in the course (for systems in equilibrium) to study the distribution of electrons in the conduction band, and that of holes in the valence band. Consider that, inside a given band, the carriers are in equilibrium among themselves, with a speciﬁc chemical potential for each band. This is justiﬁed by the fact that the phenomena responsible for transitions between conduction states, or between valence states, are much faster than the recombination of electronhole pairs. One now assumes in sections II exactly contains one electron-hole the total number of electrons contained equal to 1 ; the same assumption is valid in the valence states.

and III that the quantum box pair (regime of weak excitation) : in all the conduction states is thus for the total number of holes present

II. Statistical Study of the Occupation Factor of the Conduction Band States II.1. Write the Fermi distribution for a given electronic state |n, l, s versus β = 1/kB T and µ, the chemical potential of the conduction electrons. Show that this distribution factor fn,l,s only depends of the index n ; it will be written fn in the following. II.2. Expressing that the total number of electrons is 1, ﬁnd a relation between µ and β. II.3. Show that µ < Ec . II.4. Show that fn is a decreasing function of n. Deduce that fn ≤

1 (n + 1)(n + 2)

(Hint : give an upper bound and a lower bound for the population of electronic states of energy smaller or equal to that of level n). II.5. Deduce that for states other than the fundamental one, it is very reasonable to approximate the Fermi-Dirac distribution function by a MaxwellBoltzmann distribution and this independently of temperature. Deduce that the chemical potential is deﬁned by the following implicit relation : 1 = g(β, w) , with 2

w = eβ(µ−Ec )

and g(β, w) =

w +w (n + 1)e−βnωc w+1 n≥1

(1)

Problem 2001 : Quantum Boxes and Optoelectronics

241

II.6. Show that the function g deﬁned in II.5 is an increasing function of w and a decreasing function of β. Deduce that the system fugacity w decreases when the temperature increases and that µ is a decreasing function of T . II.7. Specify the chemical potential limits at high and low temperatures. II.8. Deduce that beyond a critical temperature Tc – to be estimated only in the next questions – it becomes legitimate to make a Maxwell-Boltzmann-type approximation for the occupation factor of the fundamental level. II.9. Now one tries to estimate Tc . Show that expression (1) can also be written under the form : w w 1 = −w+ 2 w+1 (1 − e−βωc )2 [Hint : as a preliminary, calculate the sum χ(β) =

(2) e−βnωc and compare

n≥1

it to the sum entering into equation (1)]. II.10. Consider that the Maxwell-Boltzmann approximation is valid for the 1 fundamental state if f0 ≤ . Hence deduce Tc versus ωc . e+1 II.11. For a temperature higher than Tc show that : fn = e−βnωc

(1 − e−βωc )2 2

(3)

III. Statistics of holes III.1. Give the average number of electrons in a valence band state |n , l , s of energy En , versus the chemical potential µv of the valence electrons. (Recall : µv is diﬀerent from µ because the quantum box is out of equilibrium when an electron-hole pair is injected into it.) III.2. Deduce the average number of holes hn ,l ,s present in this valence state of energy En . Show that the holes follow a Fermi-Dirac statistics, if one associates the energy −En to this ﬁctitious particle. Write the hole chemical potential versus µv . It will thus be possible to easily adapt the results of section II, obtained for the conduction electrons, to the statistical study of the holes in the valence band.

242

Exercises and Problems

IV. Radiative lifetime of the conﬁned carriers Consider a quantum box containing an electron-hole pair at the instant t = 0 ; several physical phenomena can produce its return to equilibrium. In the present section IV we only consider the radiative recombination of the charge carriers through spontaneous emission : the deexcitation of the electron from the conduction band to the valence band generates the emission of a photon. However an optical transition between a conduction state |n, l, s and a valence state |n , l , s is only allowed if speciﬁc selection rules are satisﬁed. For the InAs quantum boxes, one can consider, to a very good approximation, that for an allowed transition n = n , l = l and s = s and that each allowed optical transition has the same strength. Consequently, the probability per unit time of radiative recombination for the electron-hole pair is simply given by fn,l,s · hn,l,s 1 = τr τ0

(4)

n,l,s

The time τr is called the radiative lifetime of the electron-hole pair ; the time τ0 is a constant characteristic of the strength of the allowed optical transitions.

nsec

IV.1. Show that, if only radiative recombinations take place and if the box contains one electron-hole pair at t = 0, the average number of electrons in the conduction band can be written e−t/τr for t > 0.

Fig. 4: Lifetime τ and radiative yield η measured for InAs quantum boxes ; the curves are the result of the theoretical model developed in the present problem. The scales of τ, η and T are logarithmic. IV.2. Now the temperature is zero. Calculate the radiative lifetime using (4). The lifetime τ of the electron-hole pair can be measured. Its experimental variation versus T is shown in Fig. 4.

Problem 2001 : Quantum Boxes and Optoelectronics

243

IV.3. Assume that at low temperature one can identify the lifetime τ and the radiative lifetime τr . Deduce τ0 from the data of Fig. 4. IV.4. From general arguments show that the radiative lifetime τr is longer for T = 0 than for T = 0. (Do not try to calculate τr in this question.) Now the temperature is in the 100–300K range. For InAs quantum boxes, some studies show that the energy spacing ω between quantum levels is of the order of 15 meV for the valence band and of 100 meV for the conduction band. IV.5. Estimate Tc for the conduction electrons and the valence holes using the result of II.10. Show that a zero temperature approximation is valid at room temperature for the distribution function of the conduction states. What is the approximation then valid for the valence states ? IV.6. Deduce the dependence of the radiative lifetime versus temperature T for T > Tc . IV.7. Compare to the experimental result for τ shown in Fig. 4. Which variation is explained by our model ? What part of the variation remains unexplained ? IV.8. From the experimental values of τ at T = 200 K and at T = 0 K deduce an estimate of the energy spacing between hole levels. Compare to the value of 15 meV suggested in IV.4.

V. Non-Radiative Recombinations

trap trap

Fig. 5 : A simple two-step mechanism of non-radiative recombination. The electron-hole pair can also recombine without emission of a photon (nonradiative recombination). A simple two-step mechanism of non-radiative recombination is schematized as follows : (a) a trap located in the neighborhood of the quantum box captures the electron ; (b) the electron then recombines with

244

Exercises and Problems

the hole, without emission of a photon. The symmetrical mechanism relying on the initial capture of a hole is also possible. One expects that such a phenomenon will be the more likely as the captured electron (or hole) lies in an excited state of the quantum box, because of its more extended wave function of larger amplitude at the trap location. The radiative and non-radiative recombinations are mutually independent. Write 1/τnr for the probability per unit time of non-radiative recombination of the electron-hole pair ; the total recombination probability per unit time is equal to 1/τ . The dependence of the emission quantum yield η of the InAs quantum boxes is plotted versus T in the Fig. 4 above : η is deﬁned as the fraction of recombinations occuring with emission of a photon. V.1. Consider a quantum box containing an electron-hole pair at the instant t = 0. Write the diﬀerential equation describing the evolution versus time of the average number of electrons in the conduction band, assuming that the electron-hole pair can recombine either with or without emission of radiation. V.2. Deduce the expressions of τ and η versus τr and τnr . V.3. Comment on the experimental dependence of η versus T . Why is it consistent with the experimental dependence of τ ? Assume that the carriers, electrons or holes, can be trapped inside the nonradiative recombination center only if they lie in a level of the quantum box of large enough energy (index n higher than or equal to n0 ). Write the nonradiative recombination rate under the following form : 1 =γ τnr

(fn,l,s + tn,l,s )

(5)

n≥n0 ,l,s

V.4. From the experimental results deduce an estimate of the index n0 for the InAs quantum boxes. [Hint : work in a temperature range where the radiative recombination rate is negligible with respect to the non-radiative recombination rate and write a simple expression versus 1/T , to ﬁrst order in η, then in ln η. You can assume, and justify the validity of this approximation, that the degeneracy of any level of index higher than or equal to n0 can be replaced by 2(n0 + 1)]. This simple description of the radiative and non-radiative recombination processes provides a very satisfactory understanding of the experimental dependences of the electron-hole pairs lifetime and of the radiative yield of the InAs

Problem 2002 : Physical Foundations of Spintronics

245

quantum boxes (see Fig. 4). These experimental results and their modeling are taken from a study performed at the laboratory of Photonics and Nanostructures of the CNRS (French National Center of Scientiﬁc Research) at Bagneux, France.

Problem 2002 : Physical Foundations of Spintronics A promising research ﬁeld in material physics consists in using the electron spin observable as a medium of information (spintronics). This observable has the advantage of not being directly aﬀected by the electrostatic ﬂuctuations, contrarly to the space observables (position, momentum) of the electron. Thus, whereas in a standard semiconductor the momentum relaxation time of an electron does not exceed 10−12 sec, its spin relaxation time can reach 10−8 sec. Here we will limit ourselves to the two-dimensional motion of electrons (quantum well structure). The electron eﬀective mass is m. The accessible domain is the rectangle [0, Lx ] × [0, Ly ] of the xOy plane. The periodic limit conditions will be chosen at the edge of this rectangle. Part II can be treated if the results indicated in part I are assumed.

Part I : Quantum Mechanics

Consider the motion of an electron described by the hamiltonian ˆ2 ˆ 0 = p + α (ˆ σx pˆx + σ H ˆy pˆy ) 2m

α≥0

is the momentum operator of the electron and the σ where pˆ = −i∇ ˆi (i = x, y, z) represent the Pauli matrices. Do not try to justify this hamiltonian. Assume that at any time t the electron state is factorized into a space part and a spin part, the space part being a plane wave of wave vector k. This normalized state is thus written

eik·r eik·r [a+ (t)|+ + a− (t)|−] = Lx Ly Lx Ly

Å

ã a+ (t) a− (t)

with |a+ (t)|2 + |a− (t)|2 = 1

(1) (2)

where |± are the eigenvectors of σ ˆz . One will write kx = k cos φ, ky = k sin φ, with k ≥ 0, 0 ≤ φ < 2π.

246

Exercises and Problems

I.1. Brieﬂy justify that one can indeed search a solution of the Schroedinger equation under the form given in (2), where k and φ are time-independent. Write the evolution equations of the coeﬃcients a± (t). I.2

ã a+ (t) des(a) For ﬁxed k and φ, show that the evolution of vector a (t) Å− ã a+ (t) d = cribing the spin state can be put under the form : i dt a− (t) ã Å a+ (t) ˆs ˆ s is a 2 × 2 hermitian matrix, that will be expliH where H a− (t) citly written versus k, φ, and the constants α, and m. ˆ s are : E± = (b) Show that the two eigenenergies of the spin hamiltonian H 2 k 2 ± αk with the corresponding eigenstates : 2m ã ã Å Å 1 1 1 1 √ |χ (3) = |χ+ = √ − eiφ −eiφ 2 2 Å

(c) Deduce that the set of |Ψk,± states :

eik·r |Ψk,+ = 2LxLy

Å

1 eiφ

ã

eik·r |Ψk,− = 2Lx Ly

Å

1 −eiφ

ã (4)

with ki = 2πni /Li (i = x, y and ni positive, negative or null integers) constitute a basis, well adapted to the problem, of the space of the one-electron states. I.3 Assume that at t = 0 the state is of the type (2) and write ω = αk. (a) Decompose this initial state on the |Ψk,± basis. (b) Deduce the expression of a± (t) versus a± (0). I.4 One deﬁnes sz (t), the average value at time t of the z-component of the ˆz . Show that : electron spin Sˆz = (/2) σ

sz (t) = sz (0) cos(2ωt) + sin(2ωt) Im a+ (0)∗ a− (0) e−iφ lying along the z-axis. The I.5 The system is installed in a magnetic ﬁeld B on the orbital variables is neglected. The hamiltonian thus becomes eﬀect of B ˆ =H ˆ 0 − γB Sˆz . H (a) Why does the momentum k remain a good quantum number ? Deduce ˆ s(B) can still be deﬁned. You that, for a ﬁxed k, a spin hamiltonian H ˆ s , γ, B and σ will express it versus H ˆz . (B) ˆ s are (b) Show that the eigenenergies of H (B)

E±

=

» 2 k 2 ± α2 k 2 + γ 2 B 2 /4 2m

Problem 2002 : Physical Foundations of Spintronics

247

(γB αk). The expression of the eigenvectors (c) Consider a weak ﬁeld B ˆ of Hs to ﬁrst order in B are then given by (B)

|χ+ = |χ+ −

γB |χ− 4αk

(B)

|χ− = |χ− +

γB |χ+ 4αk

(B) Deduce that the average values of Sˆz in the states |χ± are (B)

(B)

χ+ |Sˆz |χ+ = −

γB 4αk

(B)

(B)

χ− |Sˆz |χ− = +

(B)

γB 4αk

(B)

(B)

(d) What are the values of |χ± and of the matrix elements χ± |Sˆz |χ± , (such that γB αk). in the case of a strong ﬁeld B Part II : Statistical Physics In this part, we are dealing with the properties of an assembly of N electrons, assumed to be without any mutual interaction, each of them being submitted to the hamiltonian studied in Part I. We will ﬁrst assume that there is no magnetic ﬁeld ; the energy levels are then distributed into two branches as obtained in I.2 : E± (k) =

2 k 2 ± αk 2m

(5)

with k = |k|. We will write k0 = mα/ and 0 = 2 k02 /(2m) = mα2 /2. II.1. We ﬁrst consider the case α = 0. Brieﬂy recall why the density of states D0 () of either branch is independent of for this two-dimensional problem. Show that D0 () = mLx Ly /(2π2 ). II.2. We return to the real problem with α > 0 and now consider the branch E+ (k). (a) Plot the dispersion law E+ (k) versus k. What is the energy range accessible in this branch ? (b) Express k versus k0 , 0 , m, and the energy for this branch. Deduce the density of states : ã Å … 0 (6) D+ () = D0 () 1 − + 0 Plot D+ ()/D0 () versus /0 . II.3. One is now interested by the branch E− (k). (a) Plot E− (k) versus k and indicate the accessible energy range. How many k’s are possible for a given energy ?

248

Exercises and Problems

(b) Show that the density of states of this branch is given by … 0 < 0 : D− () = 2D0 () + 0 Å ã … 0 > 0 : D− () = D0 () 1 + + 0

(7)

II.4. In the rest of the problem, the temperature is taken as zero. (a) Qualitatively explain how the Fermi energy changes with the electrons number N . Show that if N is small enough, only the branch E− is ﬁlled. (b) What is the value of the Fermi energy F when the E+ branch begins to be ﬁlled ? Calculate the number of electrons N ∗ at this point, versus Lx , Ly , m, and 0 . (c) Where are the wave vectors k corresponding to the ﬁlled states F < 0 located ? II.5. Assume in what follows that F > 0. (a) What are the wave vectors k for the occupied states in either branch E± ? (b) Calculate the Fermi wave vectors kF+ and kF− versus F , k0 , 0 , m and . (c) For N > N ∗ , calculate N − N ∗ and deduce the relation : Lx Ly m (20 + F ) π2 II.6. A weak magnetic ﬁeld is now applied on the sample. N=

(a) Using the result of I.5, explain why the Fermi energy does not vary to ﬁrst order in B. (b) Calculate the magnetization along z of the electrons gas, to ﬁrst order in B. II.7. Now one analyzes the possibility to maintain a magnetization in the system for some time in the absence of magnetic ﬁeld. At time t = 0, each electron is prepared in the spin state |+z . The magnetization along z is thus M (0) = N /2. For a given electron, the evolution of sz (t) in the absence of collisions was calculated in I.4. Now this evolution is modiﬁed due to the eﬀect of collisions. Because of the elastic collisions of an electron on the material defects, its k wave vector direction may be randomly modiﬁed. These collisions are modeled assuming that : – they occur at regular times : t1 = τ , t2 = 2τ ,..., tn = nτ ,... – the angle φn characterizing the k direction between the nth and the (n + 1)th collision is uniformly distributed over the range [0, 2π[ and decorrelated from the previous angles φn−1 , φn−2 , ..., and thus from the coeﬃcients a± (tn ). – the angles φn corresponding to diﬀerent particles are not correlated.

Problem 2002 : Physical Foundations of Spintronics

249

Under these assumptions, the total magnetization after many collisions can be calculated by making an average over angles for each electron. One writes s¯(t) for the average of sz (t) over the angles φn . Show that s¯(tn+1 ) = cos(2ωτ ) s¯(tn ). II.8. One assumes that ωτ 1. (a) Is there a large modiﬁcation of the average spin s¯ between two collisions ? (b) One considers large times t with respect to the interval τ between two collisions. Show that the average spin s¯(t) exponentially decreases and express the characteristic decrease time td versus ω and τ . (c) Take two materials : the ﬁrst one is a good conductor (a few collisions take place per unit time), the second one is a poor conductor (many collisions occur per unit time). In which of these materials is the average spin better “protected,” i.e., has a slower decrease ? Comment on the result. (d) For τ = 10−12 sec and ω=5 µeV, estimate the spin relaxation time td . II.9. To maintain a stationary magnetization in the population of N spins, one injects g± electrons per unit time, of respective spins sz = ±/2 into the system. For each spin value, the electrons leave the system with a probability per unit time equal to 1/tr . Besides, one admits that the relaxation mechanism studied in the previous question is equivalent to the exchange between the + and − spin populations : a + spin is transformed into a − spin with a probability per unit time equal to 1/td , and conversely a − spin is transformed into a + spin with a probability per unit time equal to 1/td. One calls f± (t) the populations of each of the spin states at time t, td is the spin relaxation time. ∗ (a) Write the rate equations on df± /dt. What are the spin populations f± in steady state ? ∗ f ∗ − f− (b) Deduce the relative spin population, that is, the quantity P = + ∗ ∗. f+ + f− In what type of material will the magnetization degrade not too fast ?

Solution of the Exercises and Problems

Electrostatic Screening : exercise 2000 Magnetic Susceptibility of a “Quasi-1D” Conductor : exercise 2001 Entropies of the HC Molecule : exercise 2002 Quantum Boxes and Optoelectronics : problem 2001 Physical Fundations of Spintronics : problem 2002

251

Exercise 2000 : Electrostatic Screening

253

Exercise 2000 : Electrostatic Screening I.1. The density N/Ω and the chemical potential µ are related by N 1 = Ω Ω

1 D(k)fF D (k)d3k = 4π 3

+∞

0

4πk 2 dk ã 2 k 2 exp β −µ +1 2me Å

I.2. In the high temperature and low density limit, the exponential in the denominator of the Fermi-Dirac distribution is large with respect to 1, so that Å 2 2 ã +∞ −β k − µ N 1 2me 4πk 2 dk e Ω 4π 3 0 Å ã N 2πme kB T 3/2 2eβµ Ω h2 I.3. The total energy in presence of a potential is written ε=

2 k 2 + V (r) 2me

At high temperature and low density

N

1 4π 3

e−βV (r) d3r

Å

+∞

4πk 2 dk e

−β

ã 2 k 2 −µ 2me

0

Ω

I.4. By deﬁnition one has

n(r)d3r

N= Ω

one thus identiﬁes n(r) to n(r) = n0 e−βV (r) with n0 =

1 4π 3

Å

+∞

4πk 2 dk e

−β

0

I.5. If βV (r) 1 for any r, n is uniform, i.e., n = n0 =

N Ω

ã 2 k 2 −µ 2me

254

Solution of the Exercises and Problems

II.1. The potential energy of an electron inside the electrostatic potential Φtot (r) is −e Φtot (r), the energy of a hole +e Φtot (r). The electron and hole densities are respectively equal to n(r) = n0 exp[−β(−e Φtot (r)] p(r) = p0 exp[−β(+e Φtot (r)] with n0 = p0 =

N . Ω

II.2. The total density of charges at r is equal to e[p(r) − n(r)]. The charge localized at the origin also enters into the Poisson equation : ∆Φtot (r) +

e[p(r) − n(r)] Qδ(r) + =0 ε0 εr ε0 εr

In the high temperature limit and to the lowest order p(r) − n(r) = n0 e−βeΦtot (r) − eβeΦtot (r) −2n0 eβΦtot (r) Whence the equation satisﬁed by Φtot (r) : ∆Φtot (r) = One writes λ2 = II.3. If

2n0 e2 β Q Φtot (r) − δ(r) ε0 εr ε0 εr

Ω ε0 εr kB T , this is the square of a length. N 2e2

N = 1022 m−3 T = 300 K, εr = 12.4, then λ = 30 nm. Ω

II.4. The eﬀect of the charge located at the origin cannot be felt beyond a distance of the order of λ : at a larger distance Φtot (r) is rapidly constant, the electron and hole densities are constant andopposite. The mobile electrons and holes have screened the charge at the origin.

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” Conductor

255

Exercise 2001 : Magnetic Susceptibility of a “Quasi-One-Dimensional” Conductor I.1. a) λ = 0 kB T γB

b) λ kB T γB |+1>

ε

ε

ε0 + λ

|+1>

|−1> |0 > B

ε0

ε0

0

|0 > B

0

|−1> Three levels in ε0 for B = 0 Comparable occupations of the three levels

At the temperature T , the upper levels are weakly occupied

c) kB T λ γB |+1> ε

ε

ε0 + λ

d) γB λ kB T |+1>

ε0 + λ |−1>

ε0

0

|0 > B

ε0

0

|0 > B |−1>

Comparable occupations of the three states at T . Only level | − 1 is populated. Little inﬂuence of λ, this case The magnetization is saturated, each is almost equivalent to a) ion carries a moment +γ

256

Solution of the Exercises and Problems

I.2. State |0, energy ε0 |+ ε0 + λ + γB |− ε0 + λ − γB For a single ion : z = e−βε0 + e−β(ε0 +λ+γB) + e−β(ε0 +λ−γB) = e−βε0 [1 + e−β(λ+γB) + e−β(λ−γB) ] For the whole system : Z = (z)N F = −kB T ln Z ¶ î ó© = −N kB T −βε0 + ln 1 + e−β(λ+γB) + e−β(λ−γB) ó î = N ε0 − N kB T ln 1 + e−β(λ+γB) + e−β(λ−γB) −βλ+βγB − e−βλ−βγB e ∂F = Nγ M =− ∂B 1 + e−β(λ+γB) + e−β(λ−γB) For B → 0, M ∼

N γ2βγBe−βλ i.e., 1 + 2e−βλ χ=

2N γ 2 βe−βλ Ω(1 + 2e−βλ )

I.3. Discussion of the limit cases : (a) λ = 0 ex − e−x with x = βγB M = Nγ 1 + ex + e−x 2N γ 2 (Curie law). In the limit x → 0, one obtains χ = 3ΩkB T (b) βλ 1 βγB = x M N γe−βλ (ex − e−x ) χ

2N γ 2 e−βλ ΩkB T

(Curie law attenuated by the thermal activation towards the magnetic level). (c) 1 βλ x M

2N γ 2 βB N γ[x − βλ + βλ + x] = 3 3 χ=

This situation is analogous to a).

2N γ 2 3ΩkB T

Exercise 2001 : Magnetic Susceptibility of a “Quasi-1D” Conductor

257

(d) βγB βλ 1 M Nγ The magnetization is saturated. I.4. The high temperature behavior of χ in the ﬁgure suggests a Curie law. The expression of χ obtained in I.2. vanishes for β → ∞ and for β → 0. It is positive and goes through a maximum for β of the order of λ1 (exactly for 1 − βλ+ 2e−βλ = 0, i.e., βλ = 1.47). From the temperature of the χ maximum λ can be deduced. Note : in fact, for the system studied in the present experiment, the ε0 + λ level contains 3 states, and not 2 like in this exercise based on a simpliﬁed model. II. z=

+S

e−lβγB = eβSγB

−S

2S

e−lβγB = eβSγB

l=0

1 − e−(2S+1)βγB 1 − e−βγB

ñ

1 − e−(2S+1)x f = −kB T ln z = −kB T SβγB + ln 1 − e−x m=−

ô with x = βγB

∂f ∂f (2S + 1)γβe−(2S+1)x kB T βγe−x = − γβ = γS + kB T − ∂B ∂x 1 − e−x 1 − e−(2S+1)x 2S + 1 1 m = S + (2S+1)x − γ e − 1 ex − 1

3 limits : x → +∞ x → −∞ x → 0 (a) x → +∞ (high positive ﬁeld, low temperature) m →S γ

saturation

(b) x → −∞ (high negative ﬁeld, low temperature) m → S − (2S + 1) + 1 = −S γ

saturation

(c) x → 0 : (low ﬁeld, high temperature). One expects m to be proportional to B, which allows to deﬁne a susceptibility. Therefore one has to develop m up to the term in x, which requires to express

258

Solution of the Exercises and Problems

the denominators to third order. 2S + 1 m S+ − 2 3 γ (2S + 1)x + (2S + 1)2 x2 + (2S + 1)3 x6 + ... x + S(S + 1) γBS(S + 1) m x = γ 3 3kB T

x2 2

1 3 + x6 + . . .

(Curie law)

Exercise 2002 : Entropies of the HC Molecule ò 5 Ω 3/2 + N kB (2πmk T ) B N h3 2 ò ï Ω 3/2 = −kB T ln (2πmk T ) B N h3 ï

1.

S3D = N kB ln µ3D 2. Z2D S2D µ2D

Å

ãN A 2πmkB T h2 ï ò A = N kB ln (2πmkB T ) + 2N kB N h2 ï ò A = −kB T ln (2πmk T ) B N h2 1 = N!

A molecule is spontaneously adsorbed if µ2D < µ3D . 3. If the adsorbed ﬁlm is mobile on the surface : ∆S = S2D − S3D ï ò Ah 1 ∆S = N kB ln − N kB 2 Ω(2πmkB T )1/2 4. For the only translation degrees of HC : S3D = 152.9 J.K−1 S2D = 48.2 J.K−1 The average 3D distance is of 3.3 nm, it is 0.11 nm between molecules on the surface (which is less than an interatomic distance, but one has also to account for the large corrugation of the surface).

Problem 2001 : Quantum Boxes and Optoelectronics

5. ZN =

259

1 (zrot )N (ztr )N N!

Ç å βp2φ dφdpφ exp − z1rot = h 2IA å Ç +∞ βp2φ 2π dpφ exp − = h −∞ 2IA φ 2π = 2πIA kB T h å Ç ã Å βp2φ dθdpθ dφdpφ βp2θ exp − z2rot = exp − h2 2I 2I sin2 θ

6.

7.

=

8π 2 IkB T h2

4π 2 kB T = 2 IkB T , for a symmetrical mo2hcB h lecule, such that its physical situation is identical after a 180˚ rotation. But HC is not symmetrical The expression in the course is

8. Rotation entropy of HC : Frot = −N kB T ln z2rot (the factor for indistinguishability is included inside the translation term). ò ï ∂Frot ∂ ln z2rot = N kB ln z2rot + T Srot = − ∂T ∂T −1 Srot = 33.3 J · K

Problem 2001 : Quantum Boxes and Optoelectronics I.1. Owing to the spin degeneracies one has nEc =

2 eβ(Ec −µ)

+1

and nEv =

2 eβ(Ev −µ)

+1

260

Solution of the Exercises and Problems

I.2. At any temperature nEc + nEv = 2, in the present approximation in which the total number of electrons is 2. It follows that 1 eβ(Ec −µ)

+1

+

1 eβ(Ev −µ)

+1

= 1,

an implicit relation between µ and β

I.3. This equation can be reexpressed, noting x = eβµ , y = eβEc and z = eβEv : x x + =1 y+x z+x that is, zx + yx + 2x2 = zy + zx + yx + x2 , x2 = yz, whence the result Ec + Ev , independently of temperature. µ= 2 I.4. One immediately deduces from I.3 nEc =

2 eβ(Ec −µ)

2

=

+1

e

β(Ec −Ev ) 2

+1

2 = 4 × 10−9 . e20 + 1

At 300 K with Ec − Ev = 1 eV, nEc =

At 300 K, at thermal equilibrium, the probability for the conduction levels to be occupied is thus extremely small. II.1. fnls =

1 eβ(Enls −µ)

+1

As the energy Enls only depends of n, consequently we will write fnls = fn . II.2. Noting gn for the degeneracy of state n, one has gn fn = 2(n + 1)fn n = 1 = n

n

an implicit relation between µ and β. II.3. If one had µ > Ec , 2f0 would be larger than 1, in contradiction with n = 1. II.4. En is an increasing function of n, as is also eβ(En−µ) , whence the result. One has f0 < 1/2, and fn+1 < fn , so that : 1≥

n i=0

gi fi >

n

gi fn = 2fn

i=0

n

(n + 1) = 2fn

i=0

i.e., fn ≤

1 (n + 1)(n + 2)

(n + 1)(n + 2) 2

Problem 2001 : Quantum Boxes and Optoelectronics

261

1 1 , f2 ≤ , etc. One then 6 12 β(E1 −µ) β(E2 −µ) > 5, in f2 e > 11, which justiﬁes deduces that in f1 one has e the use of the Maxwell-Boltzmann approximation, i.e.,

II.5. The relation obtained in II.4 implies f1 ≤

fnls =

1 eβ(Enls−µ)

+1

≈ e−β(En −µ)

for any excited state. Taking into account the validity of this approximation, the implicit relation between µ and β becomes 1 = 2f0 + 2

∞

(n + 1)e−β(En−µ)

n=1

With En = Ec +nωc , and w = e−β(Ec −µ) , this expression can be transformed into ∞ 1 1 (n + 1)e−βnωc = g(β, w) = −1 +w 2 w +1 n=1

=

∞ w +w (n + 1)e−βnωc w+1 n=1

(1)

II.6. Obviously g is a decreasing function of β at ﬁxed w. Moreover, ∞ ∂g 1 = + (n + 1)e−βnωc > 0, thus g is an increasing function ∂w (w + 1)2 n=1 of w. When T increases, β decreases. For g to remain constant [equal to 1/2 from (1)], the fugacity w must decrease when T increases. 1 ∂w ∂µ = w 2 (Ec − µ) + β . From II.3., µ < E0 = Ec ∂T kT ∂T ∂w < 0 implies and the ﬁrst term of the right member is > 0. Consequently, ∂T ∂µ < 0. ∂T

Since w = e−β(Ec −µ) ,

II.7. From equation (1), – when β → 0,

∞

(n + 1)e−βnωc tends to inﬁnity, thus the solution w must

n=1

tend to 0. Then µ tends to −∞. (One ﬁnds back the classical limit at high temperature) ; w 1 = , so that w 2 w+1 tends to 1 in this limit : the low temperature limit of µ is then E0 = Ec .

– when T → 0, the sum tends to 0 and (1) becomes

262

Solution of the Exercises and Problems

1 . At low temperature, w tends to 1 and f0 tends to 1/2, w−1 + 1 whereas at high temperature w tends to 0 and the Maxwell-Boltzmann statistics is then applicable to the 0 level. Since w is a monotonous function of T , this latter limit must be valid beyond a critical temperature Tc , so that then eβ(E0 −µ) 1 (w 1). II.8. f0 =

II.9. The function χ =

∞

e−βnωc is equal to

n=1

derivative

e−βωc 1 . Its = βωc −βω c 1−e e −1

∞ ∂χ = −ωc ne−βnωc ∂β n=1

allows to conveniently calculate the sum in the right member of (1), which is equal to : ∞

(n + 1)e−βnωc = χ −

n=1

=

1 eβωc 1 ∂χ = βωc + βωc ωc ∂β e − 1 (e − 1)2

2eβωc − 1 e2βωc 1 = −1 + = −1 + βω 2 βω (e c − 1) (e c − 1)2 (1 − e−βωc )2

whence the result (2). II.10. This condition is equivalent to w < 1/e, and at T = Tc one has w = 1/e. For w = 1/e, (2) becomes 1 e(e − 1) = ωc 2 2(e + 1) 1 − exp − kB Tc ωc 2 2(e + 1) = 1 − exp − i.e., 2 e +e+2 kB Tc 1+

which provides :

ωc 2(e + 1) = − ln 1 − kTc e2 + e + 2

Numerically kB Tc = 0.65 ωc. II.11. For T > Tc , the Maxwell-Boltzmann statistics is applicable to the 1 fundamental state and consequently to any state. f0 = −1 can be apw +1 proached by w, which allows to write (2) as : w 1 = 2 (1 − e−βωc )2

Problem 2001 : Quantum Boxes and Optoelectronics

263

It follows that fn = e−β(E0 +nωc −µ) = we−βnωc = e−βnωc

(1 − e−βωc )2 2

(3)

III.1. The average number of electrons in a state of spin s of energy E in the valence band is also the Fermi-Dirac distribution function : n =

1 eβ(E−µv )

+1

III.2. hn s = 1 − n =

eβ(E−µv ) 1 1 = −β(E−µ ) = β(−E+µ ) β(E−µ ) v v v + 1 e +1 e +1 e

hn s is indeed an expression of the Fermi-Dirac statistics, under the condition of writing Et = −E. The holes chemical potential is then given by µt = −µv . The results on the electrons in questions II.1. to II.11. apply to the holes (provided this transposition for the energy and for µ). IV.1. If all recombinations are radiative, dn n =− dt τr whence the result : n = n0 e−t/τr with n0 = 1. IV.2. At T = 0, both f0 and h0 are equal to 1/2, all the other occupation factors being equal to zero. Consequently, one has : f0,0,1 h0,0,1 + f0,0,−1 h0,0,−1 2f0 h0 1 1 = = = τr τ0 τ0 2τ0

i.e., τr = 2τ0

IV.3. One deduces from the ﬁgure τr = 2τ0 = 1 nsec, that is, τ0 = 0.5 nsec. IV.4. At any temperature one has fn,l,s hn,l,s 1 = τr τ0

(4)

n,l,s

Moreover the average numbers of electrons and holes are equal to 1. As the occupation factors are spin-independent, for a given value of s one gets : n,l

fn =

n ,l

hn =

1 2

264

One deduces :

Solution of the Exercises and Problems

n,l

f n hn ≤

n,l

fn

1 hn = 4 n ,l

This limit being reached at T = 0, one then ﬁnds that at any temperature the radiative lifetime is larger than or equal to its value at T = 0. IV.5. For the electrons ωc 100 meV, whereas ωv 15 meV for the holes. Tc is thus of the order of 780 K for the electrons and 117 K for the holes. Expression (2) shows that when kB T ω, w is of the order of 1, which indeed constitutes the zero temperature approximation. The MaxwellBoltzmann approximation is valid for any hole state if T > Tc = 117 K, from II.10. IV.6. One thus has f0 = 1/2 and all the other fn ’s are equal to zero, whereas from (3), (1 − e−βωv )2 h0 = 2 One then deduces that the only active transition is associated to the levels |001 and |00 − 1 (like at zero temperature) and that 1 2f0 h0 h0 (1 − e−βωv )2 = = = τr τ0 τ0 2τ0 IV.7. This model allows one to qualitatively understand the increase of the radiative lifetime with temperature, due to the redistribution of the hole probability of presence on the excited levels. This is indeed observed above 250 K, a temperature range in which a decrease of the lifetime is observed in the experiment. IV.8. At 200 K one approximately gets τr (T ) 1 (1 − e−βωv )2 = , which gives 2 1 βωv = − ln 1 − √ 2

=

2τr (0), that is,

i.e., a splitting between holes levels of 1.23 kB T for T = 200 K, that is, of 20 meV. 1 . This τr means that the variation of the average number of electrons per unit time, due to these radiative recombinations, is given by

V.1. The probability per unit time of radiative recombination is

n dn =− dt τr

Problem 2001 : Quantum Boxes and Optoelectronics

265

In presence of non radiative recombinations at the rate becomes :

1 , the total variation τnr

dn n n =− − dt τr τnr V.2. This allows to deﬁne the total lifetime τ by dn n n n =− − =− dt τr τnr τ i.e., 1 1 1 = + τ τr τnr and a radiative yield η equal to the ratio between the number of radiative recombinations per unit time and the total number of recombinations per unit time : η=

n τr n τ

=

τ = τr

1 τr 1 τr

+

1 τnr

=

1 r 1 + ττnr

V.3. When only radiative recombinations are present, τnr → ∞ and the radiative yield η is equal to 1, independently of the value of the radiative lifetime. Here the yield strongly decreases above 200 K, whereas it is constant for T < 150 K. This implies that a nonradiative recombination mechanism becomes active for T > 200 K. Correlatively, this is consistent with the decrease of the total lifetime τ in the same temperature range, whereas the results of question IV.6. show that the radiative lifetime should increase with temperature. V.4. We found in V.2 : η =

In this expression

1 = γ τnr

1 τr 1 τr

+

1 τnr

.

(fn,l,s + hn,l,s ). At suﬃciently high tem-

n≥n0 ,l,s

perature, around 350 K, τ ≈ 0.5 nsec whereas τr > 2 nsec. In this regime, τnr the nonradiative recombinations determine τ , and η ≈ . Besides, only the τr hole levels of index n0 > 0 are occupied, they satisfy a Maxwell-Boltzmann statistics. This allows to write, using (3), (1 − e−βωv )2 1 =γ 2(n + 1)e−βnωv τnr 2 n≥n0

266

Solution of the Exercises and Problems

Taking an orbital degeneracy equal to n0 + 1 for each excited level, one has (1 − e−βωv )2 1 2(n0 + 1)e−βn0 ωv =γ e−βnωv τnr 2 n≥0

= γ(1 − e

−βωv

)(n0 + 1)e

−βn0 ωv

Using the result of IV.6, 1 γ(1 − e−βωv )(n0 + 1)e−βn0 ωv (n0 + 1)e−βn0 ωv = = 2γτ0 −βωv )2 (1−e η (1 − e−βωv ) 2τ0

≈ 2γτ0 (n0 + 1)e−βn0 ωv In this temperature range ln(η) = constant +

n0 ωv . kT

The slope of η(T ) provides an approximation for n0 : between 200 and 350 K ã Å Å ã 1 n0 ωv 1 η200K − = ln(10) = ln η350K kB 200 350 which yields a value of n0 ranging between 5 and 6.

Problem 2002 : Physical Foundations of Spintronics Part I : Quantum Mechanics

I.1. A state having a plane wave eik·r for space part is an eigenstate of the momentum operator, with the eigenvalue k. The eﬀect of the hamiltonian on this type of state gives : ñ Å ãô ï 2 2Å ã Å −iφ ãò eik·r eik·r k a+ e a+ a− ˆ H = + kα a− a− eiφ a+ Lx Ly Lx Ly 2m One thus sees that, if the coeﬃcients a± verify the evolution equations ia˙ ± =

2 k 2 a± + kα e∓iφ a∓ 2m

the proposed eigenvector is a solution of the Schroedinger equation ˙ = H|ψ. ˆ i|ψ

Problem 2002 : Physical Foundations of Spintronics

267

I.2. (a) For ﬁxed k and φ, one is restricted to a two-dimensional problem, only on the spin variables, and the hamiltonian is equal to ã Å 2 k 2 0 e−iφ ˆ Hs = Id + kα eiφ 0 2m where Id represents the identity matrix. (b) It can be immediately veriﬁed that the two proposed vectors are eigenˆ s with the eigenenergies E± . vectors of H (c) The space of the one-electron states is the tensor product of the space associated to the electron motion in the rectangle [0, Lx] × [0, Ly ] with the spin space. The functions eik·r / Lx Ly are an orthonormal basis of the orbital space, provided that the choice of k allows this function to satisfy the periodical limit conditions : this requires ki = 2πni /Li , with ni positive, negative or null integer. As for the two vectors |χ± , they constitute a basis of the spin space. Thus the vectors |Ψk,± are an ˆ 0. orthonormal eigenbasis of the hamiltonian H I.3. (a) The development of the initial spin state on the pair of eigenstates |χ± is ã Å ã Å 1 1

a+ (0) = a+ (0) + a− (0)e−iφ a− (0) eiφ 2 Å ã 1

1 + a+ (0) − a− (0)e−iφ −eiφ 2 The initial state (orbital ⊕ spin) is thus given by |Ψ =

a+ (0) + a− (0)eiφ a+ (0) − a− (0)eiφ √ √ |Ψk,+ + |Ψk,− . 2 2

(b) One deduces that the electron spin state at time t is characterized by Å

a+ (t) a− (t)

ã = +

Å ã e−iE+ t/

1 −iφ a+ (0) + a− (0)e eiφ 2 Å ã e−iE− t/

1 −iφ a+ (0) − a− (0)e −eiφ 2

After development, one ﬁnds a± (t) = e−i

2 2

k t / 2m

a± (0) cos(ωt) − ia∓ (0)e∓iφ sin(ωt) .

268

Solution of the Exercises and Problems

I.4. A simple calculation gives sz (t) = =

|a+ (t)|2 − |a− (t)|2 2

cos(2ωt) |a+ (0)|2 − |a− (0)|2 2

+2 sin(2ωt) Im a∗+ (0)a− (0)e−iφ

One then deduces the result announced in the text. I.5. (a) The new term in the hamiltonian, arising from the magnetic ﬁeld, commutes with the momentumÅ operator. One Å can thus ã ã always look for eia+ a+ i k· r genstates of the form e , where is a two-component a− a− vector providing the spin state. The hamiltonian determining the evolution of this spin is written in the |±z basis : Å ã 2 k 2 −γB/2 kαe−iφ ˆ s − γB σ ˆ s(B) = H ˆz = Id + H kαeiφ γB/2 2 2m (b) One veriﬁes that the eigenenergies of the above matrix are those given in the text. σz |χ∓ = 1. One deduces (c) One notices that χ± |Sˆz |χ± = 0 and that χ± |ˆ the formulae given in the text. ˆ s(B) are practically equal to the (d) For a large B, the eigenstates of H ˆ eigenstates |±z of Sz . More precisely, for γ < 0 (electron), one has (B) (B) (B) |χ± ±|±z . Consequently, χ± |Sˆz |χ± ±/2. Part II : Statistical Physics

II.1. For α = 0, one has E± = 2 k 2 /(2m). The number of quantum states in a given spin state, with a wave vector of coordinates in the range (kx to kx +dkx ; ky to ky + dky ) is equal to d2 N =

Lx Ly Lx Ly dkx dky = k dk dφ 2 (2π) (2π)2

i.e., by integration over the angle φ : dN =

Lx Ly Lx Ly m kdk = d 2π 2π 2

The density of states is thus independent of the energy and equal to D0 () = Lx Ly m/(2π2 ).

Problem 2002 : Physical Foundations of Spintronics

269

II.2. (a) The dispersion law in this branch is written E+ (k) =

2 (k + k0 )2 − 0 2m

This corresponds to a portion of parabola, plotted in Fig.1. The spectrum of the allowed energies is [0, ∞[.

Dispersion relations E

E+

0

E

k1

k F+

k2

k0

k

F

Fig. 1 : Typical behavior of the dispersion laws E± (k) in either branch.

(b) For the branch = E+ (k), … k = −k0 + and thus

2m ( + 0 ) 2

ã Å … 0 Lx Ly dk D+ () = k = D0 () 1 − 2π d + 0

The corresponding curve is plotted in 2.

270

Solution of the Exercises and Problems 4

D(E)

3

2

1

0

1

0

1

2

3

4

5

6

7

8

9

10

E

Fig. 2: Densities of states D+ /D0 (lower curve) and D− /D0 (upper curve), plotted versus /0 . II.3. (a) For the branch = E− (k), E− (k) =

2 (k − k0 )2 − 0 2m

the variation of which is given in Fig.1. The spectrum of the permitted energies is [−0 , ∞]. – For −0 < ≤ 0, there are two possible wavevector moduli k for a given energy : … 2m k1,2 = k0 ± ( + 0 ) 2 – For > 0, there is a single allowed wavevector modulus : … 2m ( + 0 ) k = k0 + 2 (b) One then deduces, for > 0 Å ã … 0 D− () = D0 () 1 + + 0 and for < 0 (adding both contributions corresponding for either sign in the k1,2 formula) : … 0 D− () = 2D0 () + 0

Problem 2002 : Physical Foundations of Spintronics

271

The corresponding curve is plotted in Fig. 2. II.4. (a) Assume that the electrons are accommodated one next to the other in the available energy levels. Since the temperature is taken to be zero, the ﬁrst electron must be located on the fundamental level, the second one on the ﬁrst excited level, and so forth. Thus the ﬁrst levels to be ﬁlled are those of the E− branch, with k around k0 , the Fermi energy F being a little larger than −0 for N small. When N increases, the energy F also increases. It reaches 0 when all the states of the E− branch, of wave vectors ranging between 0 and 2k0 , are ﬁlled. For larger numbers of electrons, both branches E± are ﬁlled. (b) The E+ branch remains empty until F reaches the value F = 0. There are then 0 2Lx Ly m D− ()d = 0 N∗ = π2 −0 electrons on the E− branch. (c) In the E− branch, the ﬁlled states verify E− (k) k1 < k < k2 with : … 2m k1,2 = k0 ± (F + 0 ) 2

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