Critical state of the Anderson transition: between a metal and an insulator.
ABSTRACT Using a three-frequency one-dimensional kicked rotor experimentally realized with a cold atomic gas, we study the transport properties at the critical point of the metal-insulator Anderson transition. We accurately measure the time evolution of an initially localized wave packet and show that it displays at the critical point a scaling invariance characteristic of this second-order phase transition. The shape of the momentum distribution at the critical point is found to be in excellent agreement with the analytical form deduced from the self-consistent theory of localization.
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ABSTRACT: We study the interplay between disorder, interactions and decoherence induced by spontaneous emission process. Interactions are included in the Anderson model via a mean-field approximation, and a simple model for spontaneous emission is introduced. Numerical simulations allow us to study the effects of decoherence on different dynamical regimes. Physical pictures for the mechanisms at play are discussed and provide simple interpretations. Finally, we discuss the validity of scaling laws on the initial state width.The European Physical Journal Special Topics 09/2012; 217(1). · 1.80 Impact Factor
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ABSTRACT: We study the dynamics of a nonlinear one-dimensional disordered system from a spectral point of view. The spectral entropy and the Lyapunov exponent are extracted from the short time dynamics, and shown to give a pertinent characterization of the different dynamical regimes. The chaotic and self-trapped regimes are governed by log-normal laws whose origin is traced to the exponential shape of the eigenstates of the linear problem. These quantities satisfy scaling laws depending on the initial state and explain the system behaviour at longer times.New Journal of Physics 12/2012; 15(4). · 4.06 Impact Factor
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ABSTRACT: We realize experimentally a cold-atom system, the quasiperiodic kicked rotor, equivalent to the three-dimensional Anderson model of disordered solids where the anisotropy between the x direction and the y – z plane can be controlled by adjusting an experimentally accessible parameter. This allows us to study experimentally the disorder versus anisotropy phase diagram of the Anderson metal–insulator transition. Numerical and experimental data compare very well with each other and a theoretical analysis based on the self-consistent theory of localization correctly describes the observed behavior, illustrating the flexibility of cold-atom experiments for the study of transport phenomena in complex quantum systems.New Journal of Physics 01/2013; 15(6):065013. · 4.06 Impact Factor
arXiv:1005.1540v1 [cond-mat.dis-nn] 10 May 2010
Between a metal and an insulator: the critical state of the Anderson transition
Gabriel Lemarié,1, ∗Hans Lignier,2, †Dominique Delande,1Pascal Szriftgiser,2and Jean-Claude Garreau2
1Laboratoire Kastler Brossel, UPMC-Paris 6, ENS, CNRS; 4 Place Jussieu, F-75005 Paris, France
2Laboratoire de Physique des Lasers, Atomes et Molécules, Université Lille 1 Sciences et Technologies,
UMR CNRS 8523; F-59655 Villeneuve d’Ascq Cedex, France‡
(Dated: May 11, 2010)
Using a three-frequency one-dimensional kicked rotor experimentally realized with a cold atomic
gas, we study the transport properties at the critical point of the metal-insulator Anderson transition.
We accurately measure the time-evolution of an initially localized wavepacket and show that it
displays at the critical point a scaling invariance characteristic of this second-order phase transition.
The shape of the momentum distribution at the critical point is found to be in excellent agreement
with the analytical form deduced from self-consistent theory of localization.
PACS numbers: 03.75.-b, 05.45.Mt, 72.15.Rn
Different phase transitions observed in various fields
of physics often share similar characteristics . Of spe-
cial interest is the behavior of the system at the critical
point (for example scale invariance) and in its immedi-
ate vicinity (e.g. divergence of a characteristic length
scale). The advent of cold atom physics has offered new
possibilities of direct experimental observation of such
characteristics of quantum phase transitions. In this let-
ter, we show that the Anderson metal-insulator transi-
tion (which has only recently been observed with atomic
matter waves ) obeys scale invariance at the threshold,
defining a new state of matter between a metal and an
The Anderson transition takes place in 3-dimensional
(3D) disordered non-interacting systems in the meso-
scopic regime (where the transport is coherent). It in-
volves a metallic phase at low disorder associated with an
essentially diffusive transport, and an insulating phase at
large disorder where transport over long distance is inhib-
ited by interference effects: this is the so-called Ander-
son localization phenomenon . The Anderson transi-
tion is a second-order (continuous) phase transition: On
the insulating side, the localization characteristic length
ℓ diverges algebraically, ℓ ∝ |K − Kc|−νwhen K, the
control parameter, approaches the threshold Kc of the
transition. On the metallic side, similarly, the diffusion
constant vanishes algebraically D ∝ |K −Kc|s. The crit-
ical exponents s and ν are equal in 3D, and universal
(they do not depend on the microscopic details of the
system) . Only recently have these theoretical pre-
dictions been confirmed experimentally and the value of
ν = s unambiguously determined [2, 5, 6]: ν = 1.4 ± 0.3
is found perfectly compatible with ν = 1.57 ± 0.02 ob-
tained from numerical simulations of the 3D Anderson
model [7, 8].
The state of a disordered system is, in this context,
characterized by its transport properties. One can con-
sider the behavior at large distances and long times of
the (disorder) averaged intensity Green function (AIGF)
which determines the probability P(r,r′;t) for a parti-
cle to go from r to r′in time t . In the insulating
phase, the AIGF is a stationary function exponentially
P(r,r′;t) ∼ exp(−|r − r′|/2ℓ), [localized]
while in the metallic regime, it is a Gaussian expanding
P(r,r′;t) ∼ exp[−(r − r′)2/2Dt]. [diffusive] (2)
These two behaviors are however long time asymptotics.
Indeed, a localized AIGF is observed only for times
t ≫ tℓ, where tℓis the localization time (the time-scale
associated to localization). At the transition, tℓ∼ ℓ3di-
verges and the system becomes scale invariant. What is
the AIGF behavior at the critical point? In the following,
we show that it scales as:
P(r,r′;t) ∼ exp
−α|r − r′|3/2/t1/2?
, [critical] (3)
where α is a known (measurable) quantity. This defines
a new state, since the shape does not change with time.
Such a state of matter, intermediate between an insula-
tor and a metal at all scales, has never been directly ob-
served experimentally, although interesting results have
recently been published for ultrasound waves in the lo-
calized regime . The purpose of this letter is to report
the first experimental characterization of such a critical
state of the Anderson transition.
In cold atomic gases, it is possible to prepare the sys-
tem in a localized state and follow its evolution over
time; this constitutes an experimental measurement of
the (A)IGF [2, 11, 12], which is impossible to achieve in
state of art solid state physics. Observing the 3D An-
derson transition in configuration space with cold atoms
requires a disordered potential with a correlation length
comparable to the de Broglie wavelength  which has
not yet been achieved. We have recently shown [2, 5] that
it is nevertheless possible to observe Anderson localiza-
tion and the Anderson transition in momentum space by
Figure 1: First row: Measured AIGF of the quasiperiodic atomic kicked rotor at different times t in the (left to right) localized
(K,ε) = (6,0.1), critical (K = Kc,ε) = (8,0.38) and diffusive (K,ε) = (11,0.8) regimes. Second row: Appropriate rescalings
of the momentum by t0(localized), t1/3(critical) or t1/2(diffusive), bring the curves at different times into coincidence (the
vertical scales are also rescaled in order to preserve normalization). The shapes are different in the three regimes: exponential
localization, Eq. (1), in the localized regime, Gaussian shape, Eq. (2), in the diffusive regime, and the new “Airy shape”, Eq. (6),
at the critical point. Parameters are: ¯ k = 2.89, ω2 = 2π√7, ω3 = 2π√17. Time is measured in number of kicks and momentum
in units of 2?kL.
using a different system, the atomic kicked rotor (de-
scribed below), where the chaotic nature of the classical
motion replaces the disordered potential.
Our atom-optics system (see  for a detailed descrip-
tion) consists in a cloud of laser-cooled cesium atoms
(FWHM of the momentum distribution of 8?kL) inter-
acting with a pulsed (period T1 = 27.778 µs), far de-
tuned (∆ = −11.3 GHz) standing wave (wavenumber
kL = 7.4 × 106m−1and one way intensity I0 = 150
mW). The amplitude of the kicks is modulated with two
frequencies ω2and ω3. The Hamiltonian reads:
2+ K cosx[1 + εcos(ω2t)cos(ω3t)]
δ(t − n) ,
where time is measured in units of T1, space in units
of (2kL)−1, momentum in units of 2?kL/¯ k, with ¯ k =
LT1/M = 2.89 (M is the atom mass) playing the role
of an effective Planck constant ([x,p] = i¯ k) and K is the
average kick amplitude. The kicks are short enough (du-
ration τ = 0.8 µs) as compared to the atom dynamics
so that they can be considered as Dirac delta functions.
Decoherence processes, analyzed in detail in  are neg-
ligible for the typical duration of the experiment t ≃ 160
If ω2, ω3, π and ¯ k are incommensurate, this 1D
quasiperiodic kicked rotor has been shown to be equiva-
lent to a 3D disordered anisotropic system [6, 14, 15] and
to display an Anderson metal-insulator transition, as ev-
idenced by the fact that it belongs to the universality
class of the 3D Anderson model [6, 7], i.e. has the same
critical exponent ν. Here, the localization manifests it-
self in momentum space instead of configuration space.
We thus expect the AIGF to take simpler forms in mo-
mentum space, with expressions similar to Eqs. (1)-(3)
(simply replacing position r by momentum p). In order
to avoid confusion, we will use the notation Π(p,p′;t) for
the AIGF in momentum space. Thus, an initial momen-
tum distribution W(p,t = 0) will on average evolve at
time t to:
Π(p,p′;t) W(p′,0) dp′.
Experimentally, we are able to measure the momen-
tum distribution at the end of a pulse sequence (up
to 160 kicks), using Raman stimulated transitions (see
[16, 17] for details). The initial state W(p,t = 0) is
a thermal momentum distribution whose width is much
smaller than the width of the final distribution, and can
thus be approximated by a δ-function δ(p) in Eq. (5). As
a consequence, the final momentum distribution W(p,t)
faithfully measures the intensity Green function Π(p,t) ≡
Figure 1 shows the experimentally measured AIGF
Π(p,t) at various times in the localized, critical and dif-
fusive regimes. The three regimes obey different scaling
laws. In the localized regime (left column), the momen-
tum distribution is localized – i.e. it is time-independent
Figure 2: Numerically simulated time-evolution of an ini-
tially localized momentum distribution (log scale), for the
quasiperiodic kicked rotor at the critical point of the Ander-
son transition. The spreading follows an anomalous diffusion,
with ?p2(t)? ∝ t2/3, and the shape is preserved (i.e. scale in-
variant), being neither exponentially (as it is in the localized
regime), nor Gaussian (as it is in the diffusive regime). The
analytic prediction, Eq. (6) is shown as the red thick curve
at 1 million kicks. The agreement is excellent, without any
adjustable parameter. Parameters are those of fig. 1. Time is
measured in millions of kicks and momentum in units of two
recoil momenta 2?kL.
– at long times and thus scales as t0. In the diffusive
regime, the average kinetic energy ?p2(t)? increases lin-
early with time, so that the typical momentum scales as
t1/2; this is manifest in the broadening of the distribu-
tion with time seen in the right column. At the critical
point of the Anderson transition, we observe [2, 5], as
predicted by the one parameter scaling theory [18, 19],
an anomalous diffusion ?p2?(t) ∼ t2/3. This implies that
the typical momentum scales as t1/3leading to a slower
broadening of the distribution (middle column). If the
raw experimental data are rescaled according to these
laws (lower row in Fig. 1), i.e. plotted vs. pt0, pt−1/3and
pt−1/2in the localized, critical and diffusive regimes re-
spectively, curves taken at various times coincide, which
constitutes an experimental proof of the validity of the
scaling laws . The shapes of the distributions are dif-
ferent in the various regimes: exponential shape in the
localized regime, Gaussian shape in the diffusive regime.
The intermediate shape at the critical point is discussed
Figure 1 is a clear manifestation of the scale invari-
ance at the critical point. The anomalous diffusion is
not a transient behavior and the AIGF keeps the same
shape at the critical point. However, slightly off the crit-
ical point, the AIGF tends gradually to either a localized
or diffusive behavior, following the anomalous diffusion
only for short times. To confirm this observation of scale
Figure 3: (color online) (a) Experimental data for the rescaled
critical AIGF (see fig. 1) averaged over time (black circles
with error bars) and a fit given by Eq. (6), with ρ the only
fitting parameter. The agreement is clearly excellent. The
fitted value ρ is found compatible with ρ = Γ(2/3)Λc/3. The
residual does not significantly differ from zero [panel (b)]. Fits
by an exponentially localized (c) or a Gaussian (d) distribu-
tion show significant deviations.
invariance over a time scale larger than 160 kicks, we per-
formed numerical simulations of the critical dynamics up
to t = 106kicks. The result is shown in Fig. 2. The ad-
vantage of numerical simulations is that it is possible to
explore the tails of the momentum distributions (hidden
by noise in a real experiment). The anomalous diffusion
– with the characteristic sub-diffusive t1/3scaling – is
clearly visible. Obviously, the distribution is neither ex-
ponentially shaped (which would result in straight lines
in the logarithmic plot), nor has a Gaussian shape (a
parabola in the logarithmic plot).
The form of the critical AIGF can be deduced from
the self-consistent theory of localization . This mean-
field theory describes quantum transport in disordered
systems at large distances and for long times . It has
been shown to be relevant for the 1D periodically kicked
rotor  and correctly predicts a metal-insulator tran-
sition in three dimensions and the anomalous diffusion
at the threshold: D(ω) ∼ (−iω)1/3with ω the frequency
conjugated to time . (the 1/3 exponent is the counter-
part in the frequency domain of the anomalous diffusion
?p2?(t) ∼ t2/3in the time domain). Using this critical be-
havior, we can compute the AIGF for the quasiperiodic
kicked rotor . The details of the calculation will be
published elsewhere; we obtain:
where ρ is a parameter directly related to the criti-
cal quantity Λc = lim
Γ(2/3)Λc/3, where Γ is the Gamma function and Ai(x) is
t→∞?p2?/t2/3(see [2, 5]) via ρ =
the Airy function. This expression is used for the plot in
Fig. 2. The asymptotic form Eq. (3) comes simply from
the limiting behavior of the Airy function for large x and
is found perfectly intermediate between the exponential
(localized) and the Gaussian (diffusive) shapes.
The analytic prediction, Eq. (6), matches very well
the shape obtained from numerical simulations of the
quasiperiodically kicked rotor shown in Fig. 2. The only
noticeable difference is near p = 0, where the result of the
numerical simulation is slightly larger than the analytic
prediction. Note that this is only observed at very long
times, beyond 1000 kicks; this phenomenon is currently
under study. On the time scale of the experiment (160
kicks), this effect is invisible.
Figure 3 shows the comparison between the experimen-
tally measured critical AIGF and the analytic prediction,
Eq. (6). The only fitting parameter is the global scale
ρ (found in excellent agreement with ρ = Γ(2/3)Λc/3).
Although it is visually not obvious to distinguish the ob-
served shape from either an exponential shape or a Gaus-
sian shape, a careful fitting procedure gives a clear cut
result. The residual between the observed distribution
and the analytic prediction (6), shown in panel (b), is
consistently zero (within the error bars), while a fit with
an exponential shape, panel (c), of a Gaussian shape,
panel (d), displays significant deviations. This is fully
confirmed by a quantitative check of the quality of the
fit. The fit by the Airy function gives a χ2per degree
of freedom equal to 1.09 – i.e. perfectly acceptable –
while the exponential fit gives 4.5 per degree of freedom
and the Gaussian fit 8.8, two unacceptably large values.
This clearly shows that the self-consistent theory of lo-
calization accounts for the critical AIGF and its scale
In conclusion, we have studied experimentally the
transport at the threshold of the Anderson transition.
It obeys scale invariance, one fundamental property of
this second-order phase transition, and this defines a new
state, between an insulator and a metal.
form can be deduced from the self-consistent theory of
localization. Work is in progress to allow experimental
observations at longer times, which should allow us to
characterize the small deviations observed numerically,
whose origin could be multifractality.
The authors acknowledge Narei Martínez for her help
with the experiment. This work was partially financed
by Ministry of Higher Education and Research, Nord-
Pas de Calais Regional Council and FEDER through the
“Contrat de Projets Etat Region (CPER) 2007-2013” and
was granted access to the HPC resources of IDRIS un-
der the allocation 2009-96089 made by GENCI (Grand
Equipement National de Calcul Intensif).
∗Present address: Service de Physique de l’Etat Condensé
(CNRS URA 2464), IRAMIS/SPEC, CEA Saclay, F-
91191 Gif-sur-Yvette, France
†Present address: Laboratoire Aimé Cotton, Université
Paris-Sud, Bat. 505, Campus d’Orsay, F-91405 Orsay
 J. Cardy, Scaling and Renormalization in Statistical
Physics (Cambridge University Press, Cambridge, 1996).
 J. Chabé, G. Lemarié, B. Grémaud, D. Delande, P. Szrift-
giser, and J. C. Garreau, Phys. Rev. Lett. 101, 255702
 P. W. Anderson, Phys. Rev. 109, 1492 (1958).
 F. Wegner, Z. Phys. B25, 327 (1976).
 G. Lemarié, J. Chabé, P. Szriftgiser, J. C. Garreau,
B. Grémaud, and D. Delande, Phys. Rev. A 80, 043626
 G. Lemarié, B. Grémaud, and D. Delande, Europhys.
Lett. 87, 37007 (2009).
 K. Slevin and T. Ohtsuki, Phys. Rev. Lett. 82, 382
 A. Rodriguez et al., arXiv:1005.0515.
 D. Vollhardt and P. Wölfle, in Electronic Phase Transi-
tions, edited by W. Hanke and Y. V. Kopaev (Elsevier,
Amsterdam, 1992), pp. 1–78.
 S. Faez, A. Strybulevych, J. H. Page, A. Lagendijk, and
B. A. van Tiggelen, Phys. Rev. Lett. 103, 155703 (2009).
 F. L. Moore, J. C. Robinson, C. F. Bharucha, B. Sun-
daram, and M. G. Raizen, Phys. Rev. Lett. 75, 4598
 J. Billy, V. Josse, Z. Zuo, A. Bernard, B. Hambrecht,
P. Lugan, D. Clément, L. Sanchez-Palencia, P. Bouyer,
and A. Aspect, Nature 453, 891 (2008).
 R. Kuhn, O. Sigwarth, C. Miniatura, D. Delande, and
C. Müller, New J. Phys. 9, 161 (2007).
 G. Casati, I. Guarneri, and D. L. Shepelyansky, Phys.
Rev. Lett. 62, 345 (1989).
 D. M. Basko, M. A. Skvortsov, and V. E. Kravtsov, Phys.
Rev. Lett. 90, 096801 (2003).
 J. Ringot, Y. Lecoq, J. C. Garreau, and P. Szriftgiser,
Eur. Phys. J. D 7, 285 (1999).
 J. Ringot, P. Szriftgiser, and J. C. Garreau, Phys. Rev.
A 65, 013403 (2001).
 E. Abrahams, P. W. Anderson, D. C. Licciardello, and
T. V. Ramakrishnan, Phys. Rev. Lett. 42, 673 (1979).
 T. Ohtsuki and T. Kawarabayashi, J. Phys. Soc. Jpn. 66,
 H. Hu, A. Strybulevych, J. H. Page, S. E. Skipetrov, and
B. A. van Tiggelen, Nature Physics 4, 945 (2008).
 A. Altland, Phys. Rev. Lett. 71, 69 (1993).
 B. Shapiro, Phys. Rev. B 25, 4266 (1982).
 G. Lemarié, Ph.D. thesis, Université P. et M. Curie, Paris
 Extracting the precise location of the critical point is
not straightforward, especially because the experimental
data are limited to 160 kicks. As explained in , finite-
time-scaling makes it possible to determine it with a rea-
sonably small uncertainty, of the order of 0.3 on the value
of K. The experimentally observed value agrees very well
with the one extracted from numerical simulations with-
out any adjustable parameter.