. 17
( 22)


Performing the impurity average gives in the di¬usive limit for the weak-localization
interference term38
xt / 2 =x t/ 2

’ ’ i•(xcl ))
dt ( 4D ˜t
Dxt e ’t / 2
C t , ’t (x, x) = , (11.123)
2 2
x’t / 2 =x

where D0 is the di¬usion constant for a particle with energy F , D0 = vF „ /d.
Let us now obtain the equation satis¬ed by the Cooperon in the presence of a
time-dependent electromagnetic ¬eld. In that case we have for the interaction the
L1 (xt , xt , t) = ext · A(xt , t) ’ eφ(xt , t) .
™ ™ (11.124)
Since the coherence between time-reversed trajectories is partially upset, it is con-
venient to introduce arbitrary initial and ¬nal times, and we have for the phase
di¬erence between a pair of time-reversed paths
{S[xcl ] ’ S[xcl+tf ’t ]}
•[xcl ] =
t t ti


dt L1 (xcl , xcl , t) ’ L1 (xcl+tf ’t , xcl+tf ’t , t)
t ™t ™ ti
= (11.125)

37 An exception to this is discussed in Section 11.3.1.
38 In case the classical motion in the random potential is adequately described as the di¬usion
process, we immediately recover the result Eq. (11.93) for the return probability.
11.2. Weak localization 407

as the contributions to the phase di¬erence from L0 cancels, and we are left with
dt xcl (t)·A(xcl (t), t) + φ(xcl (ti + tf ’t), ti + tf ’t) ’ φ(xcl (t), t)
•[xcl ] = ™

’ xcl (ti + tf ’ t) · A(xcl (ti + tf ’ t), ti + tf ’ t) .
™ (11.126)

Introducing the shift in the time variable
t ≡t’T , T≡ (tf + ti ) (11.127)
we get
t f ’t i
xcl (t + T ) · A(xcl (t + T ), t + T )
•[xcl ] = ™
t i ’t f

’ xcl (T ’ t ) · A(xcl (T ’ t ), T ’ t )

’ φ(xcl (t + T ), t + t) + φ(xcl (T ’ t ), T ’ t ) . (11.128)

The electromagnetic ¬eld is assumed to have a negligible e¬ect on determining the
classical paths, and we can shift the time argument specifying the position on the
path to be symmetric about the moment in time T , and thereby rewrite the phase
di¬erence, t ≡ tf ’ ti ,
dt xcl · AT (xcl , t) ’ φ(xcl , t) ,
¯ ™t ¯¯ ¯¯
•[xcl ] = (11.129)
t t t

φT (x, t) = φ(x, T + t) ’ φ(x, T ’ t) (11.130)
AT (x, t) = A(x, T + t) + A(x, T ’ t) . (11.131)
An electric ¬eld can be represented solely by a scalar potential, and we imme-
diately conclude that only if the ¬eld is di¬erent on time-reversed trajectories can
it lead to destruction of phase coherence. In particular, an electric ¬eld constant in
time does not a¬ect the phase coherence, and thereby does not in¬‚uence the weak-
localization e¬ect.
The di¬erential equation corresponding to the path integral, Eq. (11.123), there-
fore gives for the Cooperon equation for the case of a time-dependent electromagnetic
408 11. Disordered conductors

‚ e ie
+ φT (xt , t) ’D0 ∇x ’ AT (x, t) Ct,t (x, x ) = δ(x ’ x ) δ(t ’ t ) .
When the sample is exposed to a time-independent magnetic ¬eld, we recover the
static Cooperon equation, Eq. (11.107).

11.3 Phase breaking in weak localization
The phase coherence between the amplitudes for pairs of time-reversed trajectories
is interrupted when the environment of the electron, besides the dominating random
potential, is taken into account. At nonzero temperatures, energy exchange due
to the interaction with the environment will partially upset the coherence between
time-reversed paths involved in the weak-localization phenomenon. The constructive
interference is then partially destroyed.
Quantitatively the e¬ect on weak localization by inelastic interactions with energy
transfers ”E of the order of the temperature, ”E ∼ kT , strongly inelastic processes,
can be understood by the observation that the single-particle Green™s function will
be additionally damped owing to interactions. If in addition to disorder we have an
interaction, say with phonons, the self-energy will in lowest order in the interaction
be changed according to

pE pE

’ pE pE pE pE
pE pE


and we will get an additional contribution to the imaginary part of the self-energy

mΣR = ’ ’ . (11.134)
2„ 2„in
11.3. Phase breaking in weak localization 409

Upon redoing the calculation leading to Eq. (11.22) for the case in question, we obtain
in the limit „in „

ζ(Q, ω) = 1 ’ + iω„ + D0 „ Q2 . (11.135)
This will in turn lead to the change in the Cooperon equation, ω ’ ω + i/„in , and
we get the real space Cooperon equation39
1 1
’ iω ’ D0 ∇2 + δ(x ’ x ) .
Cω (x, x ) = (11.136)
„in „
The e¬ect on weak localization of electron“electron interaction and electron“
phonon interaction have been studied in detail experimentally [88, 89], and can phe-
nomenologically be accounted for adequately by introducing a temperature-dependent
phase-breaking rate 1/„• in the Cooperon equation, describing the temporal expo-
nential decay C(t) ’ C(t) exp{’t/„• } of phase coherence. In many cases the in-
elastic scattering rate, 1/„in , is identical to the phase-breaking rate, 1/„• . This is
for example the case for electron“phonon interaction, as we shortly demonstrate.
However, one should keep in mind that the inelastic scattering rate is de¬ned as the
damping of an energy state for the case where all scattering processes are weighted
equally, irrespective of the amount of energy transfer. In a clean metal the energy
relaxation rate due to electron“phonon or electron“electron interaction is determined
by energy transfers of the order of the temperature as a consequence of the exclu-
sion principle (at temperatures below the Debye temperature).40 In Section 11.5
we shall soon learn that in a three-dimensional sample the energy relaxation rate
in a dirty metal is larger than in a clean metal owing to a strong enhancement of
the electron“electron interaction with small energy transfer. When calculating the
weak localization phase-breaking rate we must therefore pay special attention to the
low-energy electron“electron interaction. In a thin ¬lm or in the two-dimensional
case the energy relaxation rate even diverges in perturbation theory, owing to the
abundance of collisions with small energy transfer. However, the physically measur-
able phase-breaking rate does of course not su¬er such a divergence since the phase
change caused by an inelastic collision is given by the energy transfer times the re-
maining time to elapse on the trajectory. Collisions with energy transfer of the order
of (the phase-breaking rate) ω ∼ /„• or less are therefore ine¬cient for destroy-
ing the phase coherence between the amplitudes for traversing typical time-reversed
trajectories of duration the phase coherence time „• .41 In terms of diagrams this is
re¬‚ected by the fact that interaction lines can connect the upper and lower particle
lines in the Cooperon, whereas there are no such processes for the diagrammatic
representation of the inelastic scattering rate, as discussed in Section 11.5. This dis-
tinction is of importance in the case of a thin metallic ¬lm, the quasi two-dimensional
39 In the Cooperon, contributions from diagrams where besides impurity correlator lines interac-
tion lines connecting the retarded and advanced particle line also appear should be included for
consistency. However, for strongly inelastic processes these contributions are small.
40 For details see, for example, chapter 10 of reference [1].
41 A similar situation is the di¬erence between the transport and momentum relaxation time.

The transport relaxation time is the one appearing in the conductivity, re¬‚ecting that small angle
scattering is ine¬ective in degrading the current.
410 11. Disordered conductors

case, where there is an abundance of scatterings with small energy transfer due to
di¬usion-enhanced electron“electron interaction.
In the time-reversal invariant situation, the Cooperon is equal to the classical
probability that an electron at the Fermi level in time t returns to its starting point.
If coherence is disrupted by interactions, the constructive interference is partially de-
stroyed. This destruction of phase coherence results in the decay in time of coherence,
described by the factor exp{’t/„•} in the expression for the Cooperon, the probabil-
ity of not su¬ering a phase-breaking collision, described by the phase-breaking rate
1/„• . In view of the quantum interference picture of the weak localization e¬ect, we
shall also refer to „• as the wave function phase relaxation time.
A comprehensive understanding of the phase coherence length in weak local-
ization, the length scale L• ≡ D0 „• over which the electron di¬uses quantum
mechanically coherently, has been established, and this has given valuable informa-
tion about inelastic scattering processes. The phase coherence length L• is, at low
temperatures, much larger than the impurity mean free path l, explaining the slow
spatial variation of the Cooperon on the scale of the mean free path, which we have
repeatedly exploited.

11.3.1 Electron“phonon interaction
In this section we calculate the phase-breaking rate due to electron“phonon inter-
action using the simple interference picture described in the previous section.42 We
start from the one-electron Lagrangian, which is given by
mx2 ’ V (x) ’ eφ(x, t) ,
™ ™
L(x, x) = (11.137)
where V is the impurity potential, and the deformation potential is speci¬ed in terms
of the lattice displacement ¬eld, Eq. (2.72),
∇x · u(x, t) .
eφ(x, t) = (11.138)
It is important to note that the impurities move in phase with the distorted
lattice; hence the impurity potential has the form

Vimp (x ’ (Ri + u(x, t)) ,
V (x) = (11.139)

where Ri is the equilibrium position of the ith ion. The impurity scattering is thus
only elastic in the frame of reference that locally moves along with the lattice. We
therefore shift to this moving frame of reference by changing the electronic coordinate
according to x ’ x + u. The impurity scattering then becomes static on account of
generating additional terms of interaction. Expanding the Lagrangian Eq. (11.137)
in terms of the displacement, and neglecting terms of relative order m/M , such as
the term mu · v/2, the transformed Lagrangian can be written as L = L0 + L1 , where

42 We follow references [87] and [90].
11.3. Phase breaking in weak localization 411

L0 is given in Eq. (11.111), and43
L1 (xt , xt ) = mxt · (xt · ∇) u(xt , t) ’ x ∇ · u(xt , t) .
™ ™ ™ ™ (11.140)
In the last line we have used the relation n/2N0 = mvF /3, and the fact that the
magnitude of the velocity is conserved in elastic scattering. We therefore obtain for
the phase di¬erence44
1 1
dt {∇β u± (xcl , t) ’ ∇β u± (xcl , ’t)} x± xβ ’
•[xcl ] = δ±β x2
™t ™t , (11.141)
t t t

where summation over repeated Cartesian indices is implied, and we have chosen the
classical paths to satisfy the boundary condition, xcl = 0 = xcl .
’t/2 t/2
We must now average the quantum interference term as given in Eq. (11.116)
with respect to the lattice vibrations, and with respect to the random positions of
the impurities. Since the Lagrangian for the lattice vibrations is a quadratic form
in the displacement u, and the phase di¬erence •[xcl ] is linear in the displacement,
the phonon average can be computed by Wick™s theorem according to (see Exercise
4.108 on page 103)45
ei•[xt ] ph = e’ 2 •[xt ] ph .
cl cl 2
For the argument of the exponential we obtain (vt ≡ xcl )
t/2 t/2
(±)D±βγδ (xcl ’ xcl , t1 “ t2 )
•[xcl ]2 = dt1 dt2
t t1 t2
’t/2 ’t/2

1 1
— vt1 vt1 ’
vt2 vt2 ’
2 δ 2
δ±β vt1 δγδ vt2 , (11.143)
3 3
where the phonon correlator
∇β u± (x, t)∇δ uγ (0, 0)
D±βγδ (x, t) = (11.144)
is an even function of the time di¬erence t.
Concerning the average with respect to impurity positions, we will resort to an
approximation which, since the exponential function is a convex function, can be
expressed as the inequality
•=0 ’ 1 •[xcl ]2

C(t) C(t) imp e , (11.145)
ph imp

43 This result can also be obtained without introducing the moving frame of reference. By simply
Taylor-expanding Eq. (11.139) and using Newton™s equation we obtain a Lagrangian which di¬ers
from the one in Eq. (11.140) by only a total time derivative, and therefore generates the same
44 In neglecting the Jacobian of the nonlinear transformation to the moving frame, we neglect the

in¬‚uence of the lattice motion on the paths.
45 We have suppressed the hat on u indicating that the displacement is an operator with respect

to the lattice degrees of freedom (or we have envisaged treating the lattice vibrations in the path
integral formulation).
412 11. Disordered conductors

where we have introduced the notation for the impurity average
xt / 2 =x
Dxt Pt [xt ] (•[xcl ])2 ph
x’t / 2 =x
(•[xcl ])2 = . (11.146)
xt / 2 =x
ph imp
Dxt Pt [xt ]
x’t / 2 =x

The phase di¬erence Eq. (11.141) depends on the local velocity of the electron,
which is a meaningless quantity in Brownian motion.46 It is therefore necessary
when considering phase breaking due to electron“phonon interaction to consider the
time-reversed paths involved in the weak-localization quantum interference process
as realizations of Boltzmannian motion. At a given time, a Boltzmannian path is
completely speci¬ed by its position and by the direction of its velocity as discussed in
Section 7.4.1. We are dealing with the Markovian process described by the Boltzmann
propagator F (v, x, t; v , x , t ), where we now use the velocity as variable instead of
the momentum as used in Section 7.4.1. On account of the Markovian property, the
four-point correlation function required in Eq. (11.146) (the start and (identical) end
point and two intermediate points according to Eq. (11.143)) may be expressed as a
product of three conditional probabilities of the type Eq. (7.70), and we obtain
t/2 t/2
4m2 dˆ 1 dˆ 2
•[xcl ]2 = dt1 dt2 dx1 dx2
ph imp
t 2 (4π)2
’t/2 ’t/2

t t
— F (0, ; x1 , v1 , t1 )F (x1 , v1 , t1 ; x2 , v2 , t2 ) F (x2 , v2 , t2 ; 0, ’ )
2 2

1 1
±β γδ
— (±)D±βγδ (xcl ’ xcl , t1 “ t2 ) vt1 vt1 ’ vt2 vt2 ’
2 2
δ±β vt1 δγδ vt2 .
t1 t2
3 3


We use the notation that an angular average of the Boltzmann propagator F with
respect to one of its velocities is indicated by a bar. For example, we have for the
return probability

= F (x, t; x , 0) ≡
C(t) F (x, t; v , x , t ) . (11.148)

The space-dependent quantities may be expressed by Fourier integrals according to
Eq. (7.72). Since the Boltzmann propagator is retarded, F (v, x, t; v , x , t ) vanishes
for t earlier than t , we can expand the upper t1 -integration to in¬nity and the
46 The velocity entering in the Wiener measure, Eq. (7.103), is not the local velocity, but an
average of the velocity on a Boltzmannian path; recall Exercise 7.6 on page 197.
11.3. Phase breaking in weak localization 413

lower t2 -integration to minus in¬nity. Only thermally excited phonons contribute to
the destruction of phase coherence, and we conclude that D±βγδ (xcl ’ xcl , t1 “ t2 )
t1 t2
is essentially zero for |t1 ± t2 | ≥ /kT . We can therefore extend the domain of
integration to in¬nity with respect to |t1 ± t2 | provided that |t| /kT , and obtain
in the convex approximation
2m2 dkdk dωdω dˆ 1 dˆ 2

C(t) = C(t) exp
imp imp
(2π)8 (4π)2
2 C(t) imp

— F (v1 ; k, ω)F (v1 , v2 , k + k , ω + ω )D±βγδ (’k , ’ω )

F (v2 , k, ω) e’iωt ’ F (v2 , k, ω + 2ω ) e’i(ω+ω )t

1 1
±β γδ
— vt1 vt1 ’ vt2 vt2 ’
2 2
δ±β vt1 δγδ vt2 . (11.149)
3 3
We expect that the argument of the exponential above increases linearly in t for
large times. Since the classical return probability in three dimensions has the time
∝ t’3/2 (recall the form of the di¬usion propagator), the
dependence C(t) imp
integral above should not decrease faster than t’1/2 . Such a slow decrease is obtained
from the (k, ω)-integration only from the combination F (v1 ; k, ω) F (v2 ; k, ω), which
according to Eq. (7.76) features an infrared singular behavior (’iω + D0 k 2 )’2 for
small k and ω. In fact, it is just this combination that leads to a time-dependence
proportional to t’1/2 and, compared with that, all other contributions may be ne-
glected. For the important region of integration we thus have ω ω , since ω
is determined by the phonon correlator, which gives the large contribution to the
integral for the typical value ω kT . We are therefore allowed to approximate
F (v1 , v2 ; k + k , ω + ω ) by F (v1 , v2 ; k , ω ). In addition, the same arguments show
that the second term in the square bracket may be omitted. We thus obtain

e’t/„• ,
C(t) = C(t) (11.150)
imp imp

where the phase-breaking rate due to electron“phonon interaction is given by
1 dk dω dˆ 1 dˆ 2 1
vv ±β
F (v1 , v2 ; k , ω ) D±βγδ (k , ω ) v1 v1 ’ δ±β v1
= 2 (2π)4 2
„• (4π) 3

— v2 v2 ’ δ±β v2
. (11.151)
For simplicity we consider the Debye model where the lattice vibrations are spec-
i¬ed by the density ni and the mass M of the ions, and by the longitudinal cl and the
transverse ct sound velocities.47 We assume the phonons to have three-dimensional
47 The jellium model does not allow inclusion of Umklapp processes in the electron“phonon scat-
414 11. Disordered conductors

character. In case of longitudinal vibrations, we have the normal mode expansion of
the displacement ¬eld
u(r, t) = √ ˆ
k Qk (t) eik·r , (11.152)
N k

where N is the number of ions in the normalization volume. For the phonon average
we have

H(ωk ) cos ωk (t ’ t ) ,
Qk (t) Qk (t ) = δk,’k (11.153)
2M ωk
where ωk = cl k, provided that k is less than the cut-o¬ wave vector kD , and we
obtain for the Fourier transform of the longitudinal phonon correlator
1 ±βγδ
k k k k H(ωk ) [δ(ω ’ ωk ) + δ(ω + ωk )] .
DL (k, ω)] = (11.154)
Strictly speaking, we encounter in the above derivation H(ω) = 2n(ω) + 1, where
n is the Bose distribution function. However, the present single electron theory
does not take into account that the fermionic exclusion principle forbids scattering
of an electron into occupied states. Obedience of the Pauli exclusion principle is
incorporated by the replacement48
1 ω ω 2
H(ω) ’ coth ’ tanh = . (11.155)
2 2kT 2kT sinh kT

Upon inserting Eq. (11.154) in the expression Eq. (11.151) for the phase-breaking
rate, we encounter the directional average of expressions of the type
2 2
v 2v
k± kβ v± vβ ’ δ±β (k · v) ’ k
=k . (11.156)
3 3

Altogether the angular averages appear in the combination
§ «
⎡ ¤2
⎪ ⎪
18 ⎨ dˆ [(k · v)2 ’ k 2 v ]2 ¬
dˆ (k · v)2 ’ k 2 v ¦
2 2
v v
I(k, ω)⎣ 3 3
¦L (kl) = +
πvF k 3 ⎪ 4π ’iω + iv · k + 1/„ ⎪
4π ’iω + iv · k + 1„
© ⎭

2 kl arctan kl 3

= , (11.157)
kl ’ arctan kl kl
48 The argument is identical to the similar feature for the inelastic scattering rate or imaginary part
of the self-energy. In terms of diagrams, we recall that, in the above discussion, we have included
only the e¬ect of the kinetic or Keldysh component of the phonon propagator. Including the
retarded and advanced components makes the electron experience its fermionic nature introducing
the electron kinetic component which carries the tangent hyperbolic factor. As a consequence, a
point also elaborated in reference [91], the zero-point ¬‚uctuations of the lattice can not disrupt the
weak-localization phase coherence. A detailed discussion of the Pauli principle and the inelastic
scattering rate is given in Section 11.5 in connection with the electron“electron interaction.
11.3. Phase breaking in weak localization 415

where the result in the last line is obtained since ω = cl k vF k. For the phase-
breaking rate due to longitudinal phonons we thus obtain
1 1
dk k 2 ¦L (kl)
= . (11.158)
„•,l 6mM cl sinh cl k/kT

We note the limiting behaviors
§ 7πζ(3) (kT )3
⎪ 12 cl /l kT cl kD
⎨ nMc4
1 l
= (11.159)
⎪ π4 (kT )4
„•,l ©
30 l nMc5 kT cl /l .

The expression Eq. (11.157) for the function ¦L demonstrates in a direct way
the important compensation that takes place in the case of longitudinal phonons
between the two mechanisms contained in L1 . First, the term (k · v)2 corresponds to
mv · (v · ∇)u and represents the coupling of the electrons to the vibrating impurities.
Second, the term ’k 2 v2 /3 is connected with ’mv2 ∇ · u/3, and originates from the
interaction of the electrons with the lattice vibrations. Without this compensation,
each of the mechanisms would appear to be enhanced in an impure metal, and would
lead to an enhanced phase-breaking rate proportional to (kT )2 /(nM c3 l).
For the case of transverse vibrations, we note that DT is of similar form as
ˆ± kγ has to be replaced by (δ±γ ’ k± kγ ) and an addi-
ˆ ˆˆ
Eq. (11.154) where, however, k
tional factor of 2, which accounts for the multiplicity of transverse modes. We then
obtain a phase-breaking rate due to interaction with transverse phonons, „•,t , which
is similar to the expression in Eq. (11.158) with cl and φL replaced by ct and
3 2k 3 l3 + 3kl ’ 3(k 2 l2 + 1) arctan kl
¦T (kl) = (11.160)
k4 l4
respectively. In particular, we obtain the limiting behaviors for the phase-breaking
rate due to transverse phonons
§ 2 (kT )2
⎪ π mMc3 l ct /l kT ct kD
1 t
= (11.161)
⎪ π4 (kT )4
„•,t ©
20 l 2 mMc5 kT ct /l .

We note that in the high-temperature region, ct /l kT ct kD , the transverse
contribution is negligible in comparison with the longitudinal one if ct cl . But the
transverse rate dominates in the case where the transverse sound velocity is much
smaller than the longitudinal one. Such a situation may quite well be realized in
some amorphous metals; then, it is possible to observe a phase-breaking rate of the
form „• ∝ T 2 /l at higher, but not too high, temperatures.49 The predictions of the
theory are in good agreement with magneto-resistance measurements and carefully
conducted experiments of the temperature dependence of the resistance [92].
49 A quadratic temperature dependence of the phase-breaking rate is often observed experimen-
416 11. Disordered conductors

The physical meaning of the second term in Eq. (11.149) is as follows. It is
appreciable only if the lattice deformation stays approximately constant during the
time the electron spends on its path and leads, in this case, to a cancellation of the
¬rst term. Equivalently, electron“phonon interactions with small energy transfers do
not lead to destruction of phase coherence. The e¬ect of this term is thus e¬ectively
to introduce a lower cut-o¬ in the integral of Eq. (11.158) at wave vector k0 = 1/cl „•,l .
However, there are no realistic models of phonon spectra where this e¬ect is of
importance. We therefore have the relationships ω kT / ω 1/„• . It is
therefore no surprise that the calculated phase-breaking rates are identical to the
inelastic electron“phonon collision rates in a dirty metal [93]. When considering
phase breaking due to electron“electron interaction, which we now turn to, the small
energy transfer interactions are of importance.

11.3.2 Electron“electron interaction
In this section we consider the temperature dependence of the phase-breaking rate
due to electron“electron interaction.50 As already discussed at the beginning of this
section, special attention to electron“electron interaction with small energy transfer
must be exercised due to the di¬usion enhancement. In diagrammatic terms we
therefore need to take into account diagrams where the electron“electron interaction
connects also the upper and lower particle lines in the Cooperon.
In Section 11.5 we shall show that the e¬ective electron“electron interaction at low
energies can be represented by a ¬‚uctuating ¬eld. Its correlation function in a dirty
metal will be given by the expression in Eq. (11.269), which we henceforth employ.
We can therefore obtain the e¬ect on the Cooperon of the quasi-elastic electron“
electron interaction by averaging the Cooperon with respect to a time-dependent
electromagnetic ¬eld using the proper correlator. We therefore consider the equation
for the Cooperon in the presence of an electromagnetic ¬eld, Eq. (11.81),
‚ ie 1 1
2 ’D0 ∇x ’ AT (x, t) δ(x ’ x ) δ(t ’ t ) ,
+ Ct,t (x, x ) =
„ e’e
‚t „
where we have chosen a gauge in which the scalar potential vanishes, and 1/„ e’e is
the energy relaxation rate due to high-energy electron“electron interaction processes,
i.e. processes with energy transfers ∼ kT .51
To account for the electron“electron interaction with small energy transfers, we
must perform the Gaussian average of the Cooperon with respect to the ¬‚uctuating
¬eld. This is facilitated by writing the solution of the Cooperon equation as the path
xt =x
Dxt e’S[xt ] ,
Ct,t (R, R ) = (11.163)
xt =x

50 We follow reference [94].
51 As will become clear in the following, the separation in high- and low-energy transfers takes
place at energies of the order of the temperature. However, in the following we shall not need to
specify the separation explicitly.
11.3. Phase breaking in weak localization 417

where the Euclidean action consists of two terms
S = S0 + SA , (11.164)
™t 1
S0 [xt ] = dt1 + e’e (11.165)
4D0 „
dt1 xt1 · AT (xt1 , t1 ) .

SA [xt ] = (11.166)
In terms of diagrams, the Gaussian average corresponds to connecting the external
¬eld lines pairwise in all possible ways by the correlator of the ¬eld ¬‚uctuations,
thereby producing the e¬ect of the low-energy electron“electron interaction. Since the
¬‚uctuating vector potential appears linearly in the exponential Cooperon expression,
the Gaussian average with respect to the ¬‚uctuating ¬eld is readily done
rt =R
Drt e’( S0 [xt ] +
T SA [xt ] )
Ct,t (R, R ) = (11.167)
rt =R

where the averaged action SA is expressed in terms of the correlator of the vector
t t
dt1 dt2 xμ (t1 ) xν (t2 ) AT (xt2 , t1 )AT (xt2 , t2 ) .
SA [xt ] = ™ ™ (11.168)
μ ν
t t

If we recall the de¬nition of AT (xt , t), Eq. (11.77), we have

dq dω iq·(xt ’xt )
AT (xt2 , t1 )AT (xt2 , t2 ) = 2 e Aμ Aν
1 2

μ ν
(2π)d 2π

t1 ’ t2
t 1 + t2
— cos ω + cos ω , (11.169)
2 2
where we have introduced the notation
≡ Aμ (q, ω)Aν (’q, ’ω) .
Aμ Aν (11.170)

The electric ¬eld ¬‚uctuations could equally well have been represented by a scalar
Aμ (q, ω)Aν (’q, ’ω) Eμ (q, ω)Eν (’q, ’ω)

qμ qν
φ(q, ω)φ(’q, ’ω) .
= (11.171)
418 11. Disordered conductors

In Section 11.5 we show that the electron“electron interaction with small energy
transfers, ω kT , is determined by the temperature, T , and the conductivity of
the sample, σ0 , according to52
2kT qμ qν
Aμ Aν = . (11.172)

ω 2 σ0 q 2

Upon partial integration we notice the identity (the boundary terms are seen to
vanish as x’t = xt )
t t
ω(t1 ’ t2 )
ω(t1 + t2 )
dt1 dt2 qμ qν xμ (t1 ) xν (t2 )eiq·(xt 1 ’xt 2 ) cos
™ ™ + cos
2 2
t t

t t
ω(t1 ’ t2 )
iq·(xt 1 ’xt 2 ) ω ω(t1 + t2 )
= ’ dt1 dt2 e ’ cos
cos (11.173)
4 2 2
t t

and obtain
t t
eiq·(xt 1 ’xt 2 ) ω(t1 ’ t2 )
e2 kT dq dω ω(t1 + t2 )
[xt ] = ’ ’ cos
SA dt1 dt2 cos .
2σ0 (2π)d 2π q2 2 2
t t
Performing the integration over ω and t2 , the expression for the Cooperon becomes
xt =x t xt
™ 2 dq
q’2 (1’cos(q·(xt 1 ’x’t 1 )))
’ dt1 + „ e’e + 2eσ k T
1 (2π )2
4D 0
Dxt e
T ’t
Ct,’t (x, x ) = .
x’t =x
The singular term is regularized by remembering that in Eq. (11.174) the ω-integra-
tion actually should have been terminated, in the present context, at the large fre-
quency kT / . The factor exp{iq · (xt1 ’ xt2 )} does therefore not reduce strictly to
1 for the ¬rst term in the parenthesis in Eq. (11.174) as |xt1 ’ xt2 | ≥ (D0 /kT )1/2 ,
and this oscillating phase factor provides the convergence of the integral. We should
therefore cut o¬ the q-integral at the wave vector satisfying q = (kT / D0 )1/2 ≡ L’1 ,
as indicated by the prime on the q-integration in the two previous equations.
Introducing new variables
xt ’ x’t
xt + x’t
√ √
Rt = , rt = (11.176)
2 2
52 Since the time label T now has disappeared, no confusion should arise in the following where T
denotes the temperature. We recall Section 6.5, and note that the relation Eq. (11.172) is equivalent
to the statement that the low-frequency electron“electron interaction in a disordered conductor is
identical to the Nyquist noise in the electromagnetic ¬eld ¬‚uctuations.
11.3. Phase breaking in weak localization 419

the path integral separates in two parts53

Rt = 2R
∞ R2
t ™
’ dt t

2D 0
DRt e
Ct,’t (R, R) = dR0 0
2 2„
’∞ Rt =0 =R0

rt =0 √
t ™
2e 2 k T dq
q’2 (1’cos( 2q·rt ))
’ dt 2
t + + (2π )2
„ e’e
4D 0 σ0
— Drt e (11.177)

r0 =0

The path integral with respect to Rt gives the probability that a particle started at

position R0 at time t = 0 by di¬usion reaches the point 2 R (recall Eq. (7.103)).
Integrating this probability over all possible starting points is identical to integrating
over all ¬nal points and by normalization gives unity. We are thus left with the
expression for the Cooperon
ρt =0 t
’ dt
¯ t
1 + V (rt )
4D 0
Drt e
Ct,’t , (11.178)
2 2„
ρ0 =0

where we have introduced the notation

2e2 kT
2 dq
q ’2 1 ’ cos( 2 q · r)
V (r) = + . (11.179)
„ e’e (2π)d

As expected from translational invariance, the Cooperon is independent of position.
We have thus reduced the problem of calculating the quantum correction to the

4e D0 „
δσ(ω) = ’ dt eiωt Ct,’t (r, r) , (11.180)

in the presence of electron“electron interaction, to solving for the Green™s function
the imaginary time Schr¨dinger problem

{‚t ’ D0 δ(r ’ r ) δ(t ’ t ) .
+ V (r)} Ct,t (r, r ) = (11.181)
2 2„
In the three-dimensional case the ¬rst term in the integrand of Eq. (11.179) gives
rise to a temperature dependence of the form T 3/2 . This is the same form as the one
we shall ¬nd in Section 11.5 for the inelastic scattering rate due to electron“electron
interaction in a dirty metal. This term can thus be joined with the ¬rst term of
Eq. (11.179). We note that the description of the low-energy behavior thus joins up
smoothly with the description of the high-energy behavior, as it should.
53 This is immediately obtained by using the standard discretized representation of a path integral.
420 11. Disordered conductors

We thus have for the potential in the three-dimensional case
2 ˜
V3 (r) = + V3 (r) (11.182)
„ e’e
⎨ r LT
’e 2 r
˜3 (r) = √ kT
V (11.183)

2π 2 σ0 © L’1
r LT .
Fourier-transforming Eq. (11.181) with respect to time and taking the static limit we
{’D0 r + V3 (r)} Cω=0 (r, r ) = √ δ(r ’ r ) . (11.184)
2 2„
Solving this equation to ¬rst order in the potential V3 gives
e2 kT L ’2 2 L
C1 (0, 0, ω = 0) = ’ ’1 + Ei ’2 2
e T
4π 2 „ D0 σ0 πLT LT
where Ei is the exponential integral54 and
D0 „ e’e .
L= (11.186)
In accordance with the calculation of the inelastic lifetime in section 11.5 we have
( kT „ )1/4
∼ . (11.187)
L kF l
We can therefore expand the expression in Eq. (11.185), and obtain for the quantum
correction to the conductivity
e2 4πe2 kT L LT
δσ = 1+ ln , (11.188)
2π 2 L 2D σ L

where the second term is the correction due to collisions with small energy transfer,
proportional to T 1/4 ln T . In the two-dimensional case we obtain from Eq. (11.179)
for the potential
L’1 √
1 ’ J0 ( 2 qr)
2 e kT
V2 (r) = + dq , (11.189)
„ e’e π 2 σ0 q

where J0 denotes the Bessel function. We observe the limiting behavior of the po-
§ 2

⎪ 1 r
r LT

⎪ 4 LT

2 e kT
V2 (r) = e’e +
π 2 σ0 ⎪
„ ⎪
⎪ √ √

© 1 ’ J0 ( L2 r ) ln L2 r ’ C + ln 2 r LT ,

54 Ei(x) x
= dt for x < 0.
’∞ t
11.3. Phase breaking in weak localization 421

where C is the Euler constant.
We then get the following equation for the Cooperon in the region of large values
of r

e2 kT
2 2r 1
Cω=0 (r, r ) = √ δ(r ’ r ) . (11.191)
’D0 r + e’e + 2 ln
„ π σ0 LT 2 2„

The electron typically di¬uses coherently the distance D0 „ e’e . According to Sec-
tion 11.5, for the relaxation time in two dimensions for processes with large energy
transfers, we have
2 2
D0 N2 (0)
∼ ∼ (kF l)1/2 LT ,
„ e’e
D0 (11.192)
where N2 (0) denotes the density of states at the Fermi energy in two dimensions. The
electron thus di¬uses coherently far into the region where the potential is logarithmic,
and the slow change of the potential allows the substitution
√ √
e2 kT e2 kT 2„ e’e D0
2 2r

+ ln ln . (11.193)
„ e’e π 2 σ0 π 2 σ0
Inserting into Eq. (11.191), we can read o¬ the phase-breaking rate due to electron“
electron interaction in a dirty conductor in two dimensions55
1 kT
= ln 2π D0 N2 (0) . (11.194)
4π 2 D0 N2 (0)
The phase-breaking rate due to di¬usion-enhanced electron“electron interaction thus
depends in two dimensions linearly on the temperature at low temperatures, kT <
/„ .
The above result for the phase-breaking rate can be understood as a consequence
of the phase-breaking rate setting the lower energy cut-o¬, /„• , for the e¬ciency of
inelastic scattering events in destroying phase coherence. To show this we note that
the path integral expression for the Cooperon, Eq. (11.167), is the weighted average
with respect to di¬usive paths. Since this weight is convex, we have

e’ (•[xc l ])2
Ct ≥ Ct , (11.195)
ee imp

where the second bracket signi¬es the average with respect to di¬usive paths of the
phase di¬erence between the two interfering alternatives, Eq. (11.129),
xt / 2 =x
Dxt Pt [xt ] (•[xcl ])2 ee
x’t / 2 =x
(•[xcl ])2 = (11.196)
xt / 2 =x
ee imp
Dxt Pt [xt ]
x’t / 2 =x

55 Many experiments are performed on thin metallic ¬lms. For such a quasi-two-dimensional case
we can express the result for the phase breaking due to electron“electron interaction in a ¬lm of
e 2 kT
thickness a as „1 = 2πaσ 2 ln πaσ 0 .
• 0
422 11. Disordered conductors

and Ct is the return probability in the absence of the ¬‚uctuating ¬eld, i.e. the
denominator in the above equation. The ¬rst bracket signi¬es the Gaussian average
over the ¬‚uctuating ¬eld, i.e. the low-energy electron“electron interaction,
t/2 t/2
φ(xcl ’ xcl , t1 ’ t2 ) φ(0, 0)
(•[xcl ])2 = dt1 dt2
ee ee
t t1 t2
’t/2 ’t/2

’ φ(xcl ’ xcl , t1 + t2 ) φ(0, 0) , (11.197)
t1 t2

where we now choose to let the scalar potential represent the ¬‚uctuating ¬eld.
Fourier-transforming we encounter

dq dω iq·(xcl ’xcl )
φ(xcl ’xcl , t1 ’t2 ) φ(0, 0) =2 e φφ
t1 t2

ee imp imp
t1 t2
(2π)d 2π

— (cos ω(t1 + t2 ) ’ cos ω(t1 ’ t2 )) , (11.198)

where the correlator for the ¬‚uctuating potential is speci¬ed in Eq. (11.269). For a
di¬usion process we have, according to Eq. (7.104),56

eiq·(xt 1 ’xt 2 ) (xcl ’xcl )
cl cl
= e’D0 q |t1 ’t2 |
= eiq· imp
t1 t2

and we get
t/2 t/2
2e2 kT dq dω ’ 1 D0 q2 |t1 ’t2 |’iω(t1 ’t2 )
(•[xcl ])2 = dt1 dt2 e2 , (11.200)
ee imp
(2π)d 2π
’t/2 ’t/2

where the ω-integration is limited to the region 1/„• ¤ |ω| ¤ kT / . The averaged
phase di¬erence is seen to increase linearly in time:
1 t
(•[xcl ])2 = (11.201)
ee imp
2 „•

at a rate in accordance with the previous result for the phase-breaking rate, Eq.
The lack of e¬ectiveness in destroying phase coherence by interactions with small
energy transfers is re¬‚ected in the compensation at small frequencies between the two
cosine terms appearing in the expression for the phase di¬erence, Eq. (11.198). In the
case of di¬usion-enhanced electron“electron interaction this compensation is crucial
as there is an abundance of scattering events with small energy transfer, whereas
the compensation was immaterial for electron“phonon interaction where the typical
energy transfer is determined by the temperature.
56 The last equality is an approximation owing to the constraint, x’t/2 = xt/2 , however, for large
times a very good one.
11.4. Anomalous magneto-resistance 423

Whereas the phase-breaking rate for electron“phonon interaction is model depen-
dent, i.e. material dependent, we note the interesting feature that the phase-breaking
rate for di¬usion-enhanced electron“electron interaction is universal. In two dimen-
sions we can rewrite
e2 σ0 kT
1 kF l
= ln . (11.202)
2π 2
„• 2
Phase-breaking rates in accordance with Eq. (11.194) have been extracted from
numerous magneto-resistance measurements; see, for example, references [88] and
[89]. We note that at su¬ciently low temperatures the electron“electron interaction
dominates the phase-breaking rate in comparison with the electron“phonon interac-

11.4 Anomalous magneto-resistance
From an experimental point of view, the disruption of coherence between time-
reversed trajectories by an externally controlled magnetic ¬eld is the tool by which
to study the weak-localization e¬ect. Magneto-resistance measurements in the weak-
localization regime has considerably enhanced the available information regarding
inelastic scattering times (and spin-¬‚ip and spin-orbit scattering times). The weak-
localization e¬ect thus plays an important diagnostic role in materials science.
The in¬‚uence of a magnetic ¬eld on the Cooperon was established in Section
11.2.4, and we have the Cooperon equation
2ie 1
’iω ’ D0 {∇x ’ δ(x ’ x ) .
A(x)}2 + 1/„• Cω (x, x ) = (11.203)

We can now safely study the d.c. conductivity, i.e. assume that the external electric
¬eld is static, so that its frequency is equal to zero, ω = 0, as the Cooperon in an
external magnetic ¬eld is no longer infrared divergent. The Cooperon is formally
identical to the imaginary-time Schr¨dinger Green™s function for a ¬ctitious particle
with mass equal to /2D0 and charge 2e moving in a magnetic ¬eld (see Exercise C.1
on page 515). To solve the Cooperon equation for the magnetic ¬eld case, we can
thus refer to the equivalent quantum mechanical problem of a particle in an external
homogeneous magnetic ¬eld. Considering the case of a homogeneous magnetic ¬eld,57
and choosing the z-direction along the magnetic ¬eld and representing the vector
potential in the Landau gauge, A = B (’y, 0, 0), the corresponding Hamiltonian is
D0 D0
(ˆx + 2eB y)2 + p2 + p2 .
H= p ˆ ˆy ˆz (11.204)

The problem separates
i i
px x pz z
ψ(x, y) = e e χ(y) , (11.205)
where the function χ satis¬es the equation
D0 d2 χ(y) 1 px 2
’ ωc y ’
+ χ(y) = E χ(y) (11.206)
2 dy 2 2D0 2eB
57 The case of an inhomogeneous magnetic ¬eld is treated in reference [95].
424 11. Disordered conductors

the shifted harmonic oscillator problem where ωc is the cyclotron frequency for the
¬ctitious particle, ωc ≡ 4D0 |e|B/ , so that the energy spectrum is E = E + D0 Q2 =
˜ z
ωc (n + 1/2) + D0 Qz , n = 0, 1, 2, ...; Qz = 2πnz /Lz , nz = 0, ±1, ±2, ... . In the
particle in a magnetic ¬eld analogy, n is the orbital quantum number and px is the
quantum number describing the position of the cyclotron orbit, and describes here the
possible locations of closed loops. The Cooperon in the presence of a homogeneous
magnetic ¬eld of strength B thus has the spectral representation

ψn,px (x) ψn,px (x )
C0 (x, x ) = , (11.207)
2π 4D0 |e|B„ ’1 (n + 1/2) + D0 „ Q2 + „ /„•

where the ψn,px are the Landau wave functions
1 i
ψn,px (x) = √ e px x eiQz z χn (y ’ px /2eB) (11.208)
and χn (y) is the harmonic oscillator wave function. In accordance with the derivation
of the Cooperon equation, we can describe variations only on length scales larger than
the mean free path. The sum over the orbital quantum number n should therefore
terminate when D0 „ |e|Bnmax ∼ , i.e. at values of the order of nmax lB /l2 , where

lB ≡ ( /|e|B)1/2 is the magnetic length.
To calculate the Cooperon for equal spatial values, C0 (x, x), we actually do not
need all the information contained in Eq. (11.207), since by normalization of the wave
functions in the completeness relation we have
∞ ∞
dpx — px px 2eB 2eB
χn y ’ y’ =’ dy |χn (y)|2 = ’
’∞ 2π 2eB 2eB 2π 2π
and thereby
2eB 1
C0 (x, x) = ’ . (11.210)
4D0 |e|B„ ’1 (n + 1/2) + D0 „ Q2 + „ /„•
2π z

11.4.1 Magneto-resistance in thin ¬lms
We now consider the magneto-resistance of a ¬lm of thickness a, choosing the di-
rection of the magnetic ¬eld perpendicular to the ¬lm.58 Provided the thickness
of the ¬lm is smaller than the phase coherence length, a L• (the thin ¬lm, or
quasi-two-dimensional criterion), or the usually much weaker restriction that it is
smaller than the magnetic length, a lB , only the smallest value of Qz = 2π n/Lz ,
n = 0, ±1, ±2, ... contributes to the sum. Since the smallest value is Qz = 0, we
obtain, according to Eq. (11.59), for the quantum correction to the conductivity
e3 BD0 „ 1
δσ(B) = . (11.211)
4D0 |e|B„ ’1 (n + 1/2) + „ /„•
π2 2 a n=0
58 Thestrictly two-dimensional case can also be realized experimentally, for example by using the
two-dimensional electron gas accumulating in the inversion layer in a MOSFET or heterostructure.
11.4. Anomalous magneto-resistance 425

Employing the property of the di-gamma function ψ (see, for example, reference [96])
ψ(x + n) = ψ(x) + (11.212)

we get for the magneto-conductance

e2 ˜ ’1
f2 (4D0 |e|B
δG±β (B) = „ • ) δ±β , (11.213)
4π 2
1 1 3 1
f2 (x) = ψ + +ψ + nmax + . (11.214)
2 x 2 x
The magneto-conductance of a thin ¬lm is now obtained by subtracting the zero ¬eld
conductance. In the limit B ’ 0, the sum can be estimated to become
1 ’1
’ ln(nmax 4D0 |e|B „• ) . (11.215)
’1 (n
4D0 |e|B„ + 1/2) + „ /„•

Using the property of the di-gamma function

3 1
lim ψ +n+ ln n (11.216)
2 x

we ¬nally arrive at the low-¬eld magneto-conductance of a thin ¬lm

”G±β (B) ≡ δG±β (B) ’ δG±β (B ’ 0) = f2 (B/B• ) δ±β , (11.217)
2π 2
f2 (x) = ln x + ψ + (11.218)
and B• = /4D0|e|„ • , the (temperature-dependent) characteristic scale of the mag-
netic ¬eld for the weak-localization e¬ect, is determined by the inelastic scatter-
ing. This scale is indeed small compared with the scale for classical magneto-
resistance e¬ects Bcl ∼ m/|e|„ , as B• ∼ Bcl / F „ • .59 The weak-localization
magneto-conductance is seen to be sensitive to very small magnetic ¬elds, namely
when the magnetic length becomes comparable to the phase coherence length, lB ∼
L• , or equivalently, ωc „ ∼ / F „• . Since the impurity mean free time, „ , can be
much smaller than the phase coherence time „• , the above description can be valid
over a wide magnetic ¬eld range where classical magneto-conductance e¬ects are ab-
sent. Classical magneto-conductance e¬ects are governed by the orbit bending scale,
59 In terms of the mass of the electron we have for the mass of the ¬ctitious particle /2D0 ∼
m / F „ , and the low magnetic ¬eld sensitivity can be viewed as the result of the smallness of the
¬ctitious mass in the problem.
426 11. Disordered conductors

ωc „ ∼ 1, whereas the weak-localization quantum e¬ect sets in when a loop of typical
area L2 encloses a ¬‚ux quantum.60 We note the limiting behavior of the function

⎨ for x 1
f2 (x) = (11.219)
ln x for x 1.

The magneto-conductance is positive, and seen to have a quadratic upturn at low
¬elds, and saturates beyond the characteristic ¬eld in a universal fashion, i.e. in-
dependent of sample parameters.61 The magneto-resistance is therefore negative,
”R = ’”G/G2 , which is a distinct sign that the e¬ect is not classical, since we are
considering a macroscopic system.62
Weak localization magneto-conductance is also relevant for a three-dimensional
sample, and cleared up a long-standing mystery in the ¬eld of magneto-transport in
doped semiconductors. For details on the three-dimensional case we refer the reader
to chapter 11 of reference [1].
The negative anomalous magneto-resistance can be understood qualitatively from
the simple interference picture of the weak-localization e¬ect. The presence of the
magnetic ¬eld breaks the time-reversal invariance, and upsets the otherwise identical
values of the phase factors in the amplitudes for traversing the time-reversed weak-
localization loops. The quantum interference term for a loop c is the result of the
presence of the magnetic ¬eld changed according to

Ac A— ’ |A(B=0) |2 exp
2i e
d¯ · A(¯) = |A(B=0) |2 e ¦c
, (11.220)
x x
¯ c c

where ¦c is the ¬‚ux enclosed by the loop c. The weak-localization interference term
acquires a random phase depending on the loop size, and the strength of the magnetic
¬eld, decreasing the probability of return, and thereby increasing the conductivity.
The negative contribution from each loop in the impurity ¬eld to the conductance
is modulated in accordance with the phase shift prescription for amplitudes by the
oscillatory factor, giving the expression

|A(B=0) |2 {1’cos(2π¦c /¦0 )} e’tc /„•
G(B) ’ G(O) = . (11.221)
2π 2 c

The summation is over all classical loops in the impurity ¬eld returning to within a
distance of the mean free path to a given point, and tc is the duration for traversing
60 Beyond the low-¬eld limit, ωc „ < / F „ , the expression for the magneto-conductance can not
be given in closed form, and its derivation is more involved, since we must account for the orbit
bending due to the magnetic ¬eld, the Lorentz force [97]. When the impurity mean free time „
becomes comparable to the phase coherence time „• , we are no longer in the di¬usive regime, and
a Boltzmannian description must be introduced [98].
61 Experimental observations of the low ¬eld magneto-resistance of thin metallic ¬lms are in re-

markable good agreement with the theory. The weak-localization e¬ect is thus of importance for
extracting information about inelastic scattering strengths, which is otherwise hard to come at. For
reviews of the experimental results, see references [88] and [89].
62 The classical magneto-resistance of a macroscopic sample calculated on the basis of the Boltz-

mann equation is positive.
11.4. Anomalous magneto-resistance 427

the loop c, and ¦0 is the ¬‚ux quantum ¦0 = 2π /2|e|. The sum should be performed
weighted with the probability for the realization of the loop in question, as expressed
by the brackets. The weight of loops that are longer than the phase coherence length
is suppressed, as their coherence are destroyed by inelastic scattering. In weak mag-
netic ¬elds, only the longest loops are in¬‚uenced by the phase shift due to the mag-
netic ¬eld. It is evident from Eq. (11.221) that the low ¬eld magneto-conductance
is positive and quadratic in the ¬eld.63 The continuing monotonic behavior as a
function of the magnetic ¬eld until saturation is simply a geometric property of dif-
fusion, viz. that small di¬usive loops are proli¬c. Instead of verifying this statement,
let us turn the argument around and use our physical understanding of the weak
localization e¬ect to learn about the distribution of the areas of di¬usive loops in
two dimensions. Rewriting Eq. (11.210) we have in two dimensions
∞ nmax
B 4π B D 0

’t/„ • (n+1/2)t
C0 (x, x) = dt e e . (11.222)
„ ¦0
0 n=0

For times t > „ we can let the summation run over all natural numbers and we can
sum the geometric series to obtain

e’t/„ • 2πBD0 t 1
C0 (x, x) = dt . (11.223)
sinh 2πBD0 t
4π„ D0 t ¦0
0 ¦0

The factors independent of the magnetic ¬eld are the return probability and the
dephasing factor. Representing the factors depending on the ¬eld strength, which
describes the in¬‚uence of the magnetic ¬eld on the quantum interference process, by
its cosine transform

2πBD0 t 1 SB
= dS cos ft (S) (11.224)
sinh 2πBD0 t
¦0 ¦0

and inverting gives
1 1
ft (S) = . (11.225)
4D0 t cosh2 S
4D0 t

For the weak localization contribution to the conductance we can therefore write
∞ ∞
e’t/„ •
2e2 D0 „ BS
δG(B) = ’ dt dS ft (S) cos (11.226)
π 4π„ D0 t ¦0
0 0

and we note that ft (S) is normalized, and has the interpretation of the probability
for a di¬usive loop of duration t to enclose the area S.
For the average size of a di¬usive loop of duration t we have

S dS Sft (S) = 4D0 t ln 2 , (11.227)


. 17
( 22)