From quantum to thermal topologicalsector fluctuations of strongly interacting bosons in a ring lattice
Abstract
Inspired by recent experiments on BoseEinstein condensates in ring traps, we investigate the topological properties of the phase of a onedimensional Bose field in the presence of both thermal and quantum fluctuations – the latter ones being tuned by the depth of an optical lattice applied along the ring. In the regime of large filling of the lattice, quantum Monte Carlo simulations give direct access to the full statistics of fluctuations of the Bosefield phase, and of its winding number along the ring. At zero temperature the windingnumber (or topologicalsector) fluctuations are driven by quantum phase slips localized around a Josephson link between two lattice wells, and their susceptibility is found to jump at the superfluidMott insulator transition. At finite (but low) temperature, on the other hand, the winding number fluctuations are driven by thermal activation of nearly uniform phase twists, whose activation rate is governed by the superfluid fraction. A quantumtothermal crossover in winding number fluctuations is therefore exhibited by the system, and it is characterized by a conformational change in the topologically nontrivial configurations, from localized to uniform phase twists, which can be experimentally observed in ultracold Bose gases via matterwave interference.
pacs:
64.70.Tg, 05.30.Rt, 03.75.b, 03.75.Lm, 74.81.FaKeywords: Ultracold gases, Optical lattices, SuperfluidMott insulator transition, Topological defects, Atomic circuits
1 Introduction and main results
Ultracold atoms provide a very promising platform to investigate phenomena in mesoscopic physics, dominated by quantum coherence across the entire sample, and by strong fluctuations, both of quantum and thermal origin. Recent experiments have revealed the wide potential of atomic circuits to study quantum coherent transport [8, 31, 13, 23], reproducing fundamental phenomena such as persistent currents, hysteresis, conductance quantization, etc. In particular ring traps [36, 13, 10, 12] are emerging as a fundamental platform mimicking electric coils, fundamentally sensitive to rotation or gauge fields, and they represent the simplest, topologically nontrivial geometry to investigate different topological sectors of a manybody model.
A novel element introduced by coldatom physics is the ability to study quantum transport phenomena with particles obeying bosonic statistics. In this context, a fundamental asset of coldatom experiments is the use of matterwave interferometry (or heterodyning) to reconstruct the phase pattern of the manybody bosonic state – an aspect which has been extensively used, among other things, to reveal the rapid spatial decay of firstorder coherence in quasi1d Bose gases [20, 21], or topological defects in the phase pattern in quasi2 gases across a KosterlitzThouless (KT) transition [19]. Very recent experiments have probed the interference between a quasi1d gas in a ring trap and a nearly uniform condensate [10, 12], within a geometry which allows the full reconstruction of the phase pattern along the ring [26]. In particular the ring geometry represents the simplest, nonsimply connected platform to investigate topological defects in the phase pattern, in the form of quantized vortices
(1) 
where is the phase of the Bose field, and is the winding number of the phase pattern. The winding number partitions the Hilbert space of bosonic fields on a ring into different topological sectors. Recent experiments have shown the ability to generate finite winding numbers via a thermal quench [10] or by stirring the condensate via rotation of a potential barrier along the ring [36, 6, 13, 12]; as well as the ability to faithfully reconstruct the whole phase profile along the ring and its associated winding number by matterwave heterodyning [10, 12].
We draw inspiration from these experiments to investigate phase coherence and winding number fluctuations in a regime unexplored so far in the experiments, namely that of strong correlations among the bosons. Strong correlations are most naturally induced in coldatom setups using optical lattices [7], which can destroy condensation by driving the system towards a Mott insulating state. An optical lattice can be created experimentally along the ring trap by either using LaguerreGauss beams with high angular momentum [3] or via spatial light modulators [10, 9, 17, 2]. In this way the ring is partitioned into potential wells separated by tunnel barriers, forming a chain of bosonic Josephson junctions; and the tunnel barriers can act as nucleation centers for phase twists in the Bose field.
In the presence of finite interactions, phase twists can be generated via quantum fluctuations of the local phase of the Bose field, due to the duality between phase and density, . Given the vortexquantization condition in Eq. (1), interactions can generate quantum coherent fluctuations between different winding numbers by inducing revolutions in the phase along the ring, called quantum phase slips (QPS). QPS can be regarded as the dual phenomenon to Josephson tunneling [24], as they establish coherence between Bosefield configurations with a well defined phase, dual to the coherence between Fock states established by Josephson tunneling. Due to their fundamental importance, QPS have been the subject of intense research in superconducting nanostructures, including narrow superconducting wires [5] and Josephson junction arrays [27, 25, 14], as well as cold atoms trapped in arrays of opticallattice tubes [32] and Helium trapped in nanostructures [22]. Phase slips (either of thermal or quantum origin) are responsible for the decay of supercurrents in quasionedimensional superconducting devices and confined superfluids, and their detection so far is primarily based on transport measurements. On the other hand, the realization of QPS in ringtrapped condensates clearly opens the possibility of detecting them interferometrically using the same scheme as in the recent experiments reported in Refs. [10, 12]. But how can one interferometrically distinguish a QPS from a phase slip of thermal origin (or thermal phase slip, TPS)? And under which conditions does one cross over from a regime dominated by thermal winding number fluctuations, to a regime dominated by quantum winding number fluctuations induced by QPS?
These are precisely the questions addressed by our present paper. Based on an extensive numerical study of the quantum phase model – faithfully representing the physics of optical lattices in the large filling regime – we reconstruct the quantumtothermal crossover in winding number fluctuations across the phase diagram of the system, namely from the superfluid phase at weak coupling to the Mott insulating phase at strong coupling. In particular our findings show that QPS are localized objects, occurring in an uncorrelated fashion in different segments of the ring, and the ensuing quantum winding number fluctuations obey Poissonian statistics, growing linearly with system size. The measurable interferograms manifest the localized nature of QPS in the form of line dislocations in the interference fringes, akin to those caused by a tunnel barrier [12]. Our data strongly suggest that the susceptibility of quantum windingnumber fluctuations exhibits a jump at the quantum phase transition from superfluid to Mott insulator, endowing such a transition with a precise topological nature.
On the other hand, at finite (but sufficiently low) temperatures thermal winding number fluctuations occur between the lowest energy states in each topological sector, implying that the phase twist characterizing such sectors is rather uniformly distributed along the ring. A uniform phase twist translates into measurable interferograms whose fringes form regular spirals, similar to those observed after thermal quenches [10]. The crossover from localized QPS to extended TPS turns out to be very sharp for sufficiently weak interactions among the bosons, whereas it is broad (or even poorly defined) in the strongly interacting regime. The thermal activation of windingnumber fluctuations is governed by the superfluid density at zero temperature, and hence it may serve as a way to measure this quantity, so far elusive in coldatom experiments.
Our paper is organized as follows: Sec. 2 describes the quantum phase model and the pathintegral approach used in the numerical calculations, the pathintegral picture of thermal and quantum phase slips, and the basics of their experimental detection via matterwave interferometry; Sec. 3 focuses on the thermal regime of windingnumber fluctuations and on the superfluid/Mottinsulator transition; Sec. 4 discusses the thermal onset of windingnumber fluctuations and its relationship to the superfluid fraction; the crossover between the quantum and the thermal regime, as well as a peculiar thermal revival of quantum fluctuations above the superfluid ground state, is the subject Sec. 5; finally conclusions are drawns in Sec. 6.
2 Quantum phase model and its topological excitations
2.1 From BoseHubbard to quantum phase model
A faithful description of ultracold bosons in an optical lattice is provided by the BoseHubbard model
(2) 
where , are Bose field operators defined on a ring of length with periodic boundary conditions, , is the hopping energy and the onsite repulsion. Hereafter we assume an average integer filling of the chain. Strictly working with Bose fields turns out to be impractical when one aims at investigating the quantum or thermal fluctuations of the phase pattern in the Bose field itself: indeed the phase operator is not well defined, unless one works in the largefilling regime . In this limit one may map the Bose operators to amplitude and phase operators and neglect amplitude fluctuations, namely one can take and . This mapping reduces the BoseHubbard Hamiltonian to a quantum phase (or quantum rotor) model (QPM) – a cornerstone in the study of Josephson junction arrays [16]:
(3) 
In the following we shall express all energy scales in units of , namely we introduce the reduced interaction and reduced temperature . The QPM model is perfectly suited to the study of the quantum dynamics of the Bosefield phase in a regime in which density fluctuations are irrelevant – which is at the center of our investigation. As we shall see later (Sec. 2.3) the QPM regime of large filling is also most appropriate in the perspective of extending the existing experimental setups and detection schemes for ringtrapped condensates [10, 12] to the regime of strong correlations.
The pathintegral representation of the partition function of the QPM maps the latter model onto an effective classical XY model in (1+1) dimensions [33, 30], with a spatially anisotropic Hamiltonian
(4) 
defined on a torus, where is the size of the grid in the imaginarytime dimension (), discretized into infinitesimal steps .
The exact properties of the QPM are strictly recovered in the limit, due to the TrotterLie decomposition at the heart of the pathintegral approach [30]. Moreover the mapping of the interaction term to an XY interaction in imaginary time is accurate provided that the dimensionless parameter is small [33]. Hence is the most stringent requirement for the quality of the quantumtoclassical mapping. In the following we use , which essentially guarantees convergence of all the quantities of interest towards their limit.
With held fixed, the effective XY Hamiltonian (divided by the temperature scale) is most conveniently recast in the form
(5) 
where we have introduced the coupling constants and . It is important to remark that the coupling constants are independent of temperature, and they are also held fixed for a given fixed strength of the interaction; hence varying the temperature has the unique effect of controlling the length of the extra dimension, . On the other hand, changing the interaction strength at fixed temperature has the effect of changing both the length , and the anisotropy between the spacelike () and imaginarytimelike () couplings.
The quantumtoclassical mapping of the QPM opens the path to studying the equilibrium properties of the system via Monte Carlo simulations. We perform simulations on chains of length between and , and for a broad range of values for the ratio , covering the superfluid/Mottinsulator (SF/MI) transition. We make use of a combined update scheme comprising singlespin Metropolis updates; overrelaxation moves; Metropolis updates proposing rotations of entire chains of spins in imaginary time (which essentially behave as gigantic spins in the limit ); and Wolff cluster updates [35]. In the following we shall denote with the Monte Carlo averages, calculated from the effective classical XY model of Eq. (4), to contrast it with the general statistical averages of quantum operators, denoted with .
The pathintegral Monte Carlo approach to the QPM is particularly appealing, as the computational basis diagonalizes the phase operator, and hence one has readily access to the statistics of fluctuations – both thermal and quantum (coherent)– of the phase pattern, and in particular to its winding number. Indeed each imaginarytime slice in the pathintegral representation samples a possible phase pattern, which in turn is projectively measured in an interferometric experiment.
The winding number of the phase pattern at the th imaginarytime slice is defined as
(6) 
where
(7) 
2.2 Thermal vs. quantum phase slips in the pathintegral representation
As already mentioned in the introduction, bosonic chains can support two different types of topological excitations, leading to finite windingnumber configurations, namely quantum and thermal phase slips (QPS and TPS). The two different kinds of excitations are best visualized within the pathintegral formalism, as sketched in Fig. 1.
TPS are common to a classical XY chain, namely they correspond to the appearance of a finite winding number on all imaginarytime slices – see Fig. 1(a). This requires a massive rearrangement of the pathintegral configuration, which can be achieved in a Monte Carlo simulation via Wolff clusters percolating along the imaginarytime direction, producing incoherent winding number fluctuations on all imaginarytime slices.
The pathintegral formalism offers an ideal tool to visualize QPS. Indeed the mapping of the QPM onto a dimensional XY model as in Eq. (4) allows one to identify QPSs in the spacetime phase pattern as those configurations in which the imaginarytime evolution leads to a change in winding number between different imaginarytime slices , see Fig. 1(b). Such a situation is only possible if two spacetime topological excitations, in the form of a vortexantivortex (VAV) pair, have appeared in the system. In the presence of a finite interaction among particles, QPS originating from VAV pairs displaying a finite separation along the imaginarytime axis will always exist in the ground state. Indeed, unless , the imaginarytime direction has a finite (namely, nonzero) length (at fixed ) and a finite (namely, not infinite) coupling constant along that direction: hence the effective XY model in Eq. (4) always displays nucleation of VAV pairs separating in imaginary time – and this property persists down to zero temperature. Below the KosterlitzThouless SFMI transition, the VAV pairs are bound, implying that the QPS are localized both in real space and imaginary time, namely they extend over an imaginarytime and spatial width which are not scaling with the corresponding sizes and , respectively. In this regime, the quantum winding number fluctuations are of “virtual” nature, namely the topological sector is only weakly admixed with the sectors via localized QPS. In the pathintegral image, this correspond to shortlived (in imaginary time) fluctuations towards values. Entering into the MI phase, the unbinding of VAV pairs produces extended QPS along the imaginarytime direction, which induce strong coherent quantum fluctuations of the winding number. The separation between vortex cores may now scale with the length of the imaginarytime direction and with system size; in particular the vortex cores may wind around the imaginarytime direction, inducing a global change in the winding number [15] akin to the thermal activation of a TPS, but of coherent nature.
2.3 Visualizing topological defects via matterwave heterodyning
Recent experiments in ringtrapped BoseEinstein condensates [10, 12] have shown the unique ability to reconstruct the phase pattern along the ring, making use of interference with a condensate trapped at the center of the ring in a diskshaped potential. In the following we shall imagine that such a scheme is equally applicable to the case in which an optical lattice has been created along the ring, dividing it into a sequence of (singlemode) minicondensates with filling – obtained under the condition of the opticallattice period being significantly larger than the interparticle spacing. Moreover, as suggested by our sketch in Fig. 2, we shall imagine that the optical lattice wells are oblate along the tangential direction to the ring, so that the strongest confinement remains along the radial direction.
Under the above conditions, releasing the potentials in the plane of the ring produces a quasi2 expansion of each minicondensate, as well as of the disktrapped condensate, which is much faster along the radial direction than along the polar one. As is well known, the interference pattern among “two condensates that have never seen each other” [4] is akin to that of two gaugesymmetrybreaking matter waves with a well defined phase but a random phase relationship – which can be ascribed, as a convention, to an arbitrary phase of the condensate in the disk. For sufficiently short time scales the expansion can be approximated as being purely radial, and the resulting density profile along the radial direction at a given polar angle can be obtained approximately from the farfield radial expansion of both the matter wave in the ring and the disk. After a time of flight the ring wavefunction and the disk wavefunction in the wedge centered around the angle (see Fig. 2) read
(8) 
where and are the (realvalued) Fourier transform of the transverse profiles of the groundstate wavefunction in the potential well and in the disk respectively, calculated in ; is the phase of the minicondensate in the th lattice well, corresponding to the polar angle ; is the ring radius; and is the mass of the particles. The interference term then takes the form
(9) 
where and .
Putting together the different interference patterns developing along the radial direction results in a global pattern equivalent to the one partially sketched in Fig. 2, in which the fringe spacing (controlled by the vector) is the same at all polar angles , but the fringe pattern along the radial direction is shifted by the phase profile along the ring. This direct correspondence between the Bosefield phase and the spatial phase of the fringe pattern allows the full reconstruction of the phase structure, as done in Refs. [10, 12].
In the following sections we shall plot the interference term of Eq. (9) – without the modulation due to the Fourier transforms of the disk and ring transverse modes – to illustrate the contrasting experimental signatures of TPS and QPS.
3 Quantum regime: quantum phase slips
3.1 Superfluidinsulator transition
We begin the description of our results the from groundstate behavior at . Finitetemperature effects are essentially removed from the MonteCarlo simulations below a sizedependent temperature , which guarantees that the gapless phononlike excitations in the superfluid phase are not thermally excited. Fig. 3 shows the determination of the SF/MI transition at from the Monte Carlo calculation of the population in the momentum distribution
(10) 
which is expected to scale as at the KT quantum critical point separating the SF from the MI. The crossing point of the different curves for gives an estimate of for the critical point of the SF/MI transition. This estimate is in very good agreement with the prediction of [11], giving for the BoseHubbard model in the limit of very large, integer filling. At the same time the transition is marked by the vanishing of the superfluid fraction
(11) 
where is the free energy of the system, and is a phase twist applied in the space direction. Moreover we have denoted by
(12) 
the average Josephson coupling term, and by
(13) 
the current susceptibility. The identities between the general definition and the Monte Carlo estimator of the above quantities are strictly recovered in the limit. Fig. 3 shows that the SF/MI transition is marked by a strong suppression of the superfluid density calculated on a finite size, alluding to the presence of a jump expected for this quantity in the thermodynamic limit.
3.2 Windingnumber fluctuations and quantum phase slips
The windingnumber fluctuations at are uniquely induced by QPS, which, as seen in the pathintegral picture in Sec. 2, are related to the nucleation of VAV pairs separating in the imaginarytime dimension. Figs. 4 and 5 show the reconstructed interference pattern from representative configurations with at fixed , produced by the QMC simulation, and possessing a finite winding number ( and 2). In the SF regime of strong VAV binding, the cores of the VAV pairs remain very close to each other, thereby producing localized QPS in the phase configurations with ; such localized QPS produce very sharp fringe dislocations in the interference patterns for weak interactions , gradually spreading over an increasingly large range when increases to cross the SF/MI transition. The localized fringe dislocations reorganize the fringe pattern into a spiral, similarly to what is observed in the case of a ring with a tunnel barrier in Refs. [12, 26] acting as a phase slip center; in our case of a Josephson junction chain, tunnel barriers separate each pair of sites, and QPS spontaneously nucleate at any position in the lattice.
3.2.1 Superfluid phase
The localized nature of the QPS, and their diluteness in the confined regime , makes them statistically independent: every tunnel barrier has a finite probability of nucleating a QPS with an arbitrary sign , and QPS sum up independently to produce a total winding number which, to a first approximation, is a Poissonian process with zero average and fluctuations
(14) 
Such a linear scaling is indeed verified in the SF regime, as shown in Fig. 6(a). A more refined analysis implies the necessity to consider that each configuration may display a number of QPS which follows a Poissonian distribution with parameter ; if is the number of QPS, each of them may display phase slip at random, resulting in a which is the displacement of a symmetric random walk with steps. The resulting probability is then that of the displacement of steps of a random walk of stochastic length , namely
(15) 
This Ansatz is indeed well verified by our QMC data, as shown in Fig. 7. Fits to Eq. (15) provide estimates for the probability which are in very good agreement with those drawn from the scaling plot in Fig. 6(a).
The probability to nucleate a QPS on a given tunnel barrier can be estimated by analyzing the Hamiltonian of a single Josephson junction (JJ), namely
(16) 
where is the phase difference across the junction, is the Josephson energy, and the charging energy. Following Ref. [24] one can approximate as the (unbounded) position of a particle, so that the Hamiltonian of a JJ in Eq. (16) is the same as that of a fictitious particle in a cosine potential; in the limit (or ) the lowest band has a width given by
(17) 
associated with an effective tunneling amplitude .
The appearance of a localized QPS can be attributed to the tunneling of the variable between two minima of the cosine potential – in fact it involves the appearance of a phase difference across a junction, namely the climbing of the variable to the top of the barrier between two minima. Hence it is reasonable to expect that the QPS nucleation probability follows a similar functional dependence on as in Eq. (17). A similar result can be obtained within a WentzelKramersBrillouin approximation [29] to the tunneling problem: as plays the role of the inverse mass of the fictitious particle with position , the amplitude of the wavefunction within the barrier scales as . This exponential dependence of QPS probability on is clearly exhibited in Fig. 6(b), showing that for the windingnumber fluctuations follow the behavior with a constant. The essential singularity displayed by the quantum windingnumber fluctuations as a function of in the limit of weak interactions clearly shows that this effect is due to highly nonlinear quantum fluctuations, and it cannot be captured by a harmonic treatment such as Bogoliubov theory or Luttingerliquid theory.
3.2.2 Quantum phase transition and Mott insulator phase
As shown in Fig. 6(a), upon crossing the SFMI transition one observes a strong enhancement of the winding number fluctuations, accompanied by a breakdown of the simple linear scaling with system size (Eq. (14)). A similar behavior is also to be found in the (equaltime) current fluctuations (see Fig. 8(a)):
(18) 
Our data do not allow us to conclude about a singular behavior (such as a jump) of or across the transition. As a matter of fact, a numerical study of the thermal KT transition in the uniform 2 XY model does not show any jump singularity of the corresponding observables, namely the winding number fluctuations or the current fluctuations along 1 cuts [28]. On the other hand the a critical jump is expected for the current susceptibility of Eq. (13), given that the jump in the superfluid density in Eq. (11) at the SF/MI transition comes indeed from the current fluctuations and not from the Josephson term. Nonetheless, as is well known [34] and clearly documented by Fig. 3(b)), the numerical reconstruction of the jump is strongly limited by finitesize effects. Indeed one expects that [34]
(19) 
where is the correlation length – exponentially divergent as – and and are constants. The exponentially decreasing term in Eq. (19) introduces the first correction to linear scaling (which becomes logarithmic exactly at the transition [34], where ), and its negative sign justifies the fact that the convergence of current fluctuations in the thermodynamic limit is attained from below.
In the context of topological sector fluctuations, the counterpart to the current susceptibility is represented by the winding susceptibility
(20) 
Fig. 8(b) shows that sizable deviations from linear scaling build up in at the transition, in a very similar manner to what happens to the current susceptibility. Therefore we conclude that a jump singularity is to be expected in the winding susceptibility, and that may obey a similar scaling as for in Eq. (19). Indeed strong windingnumber correlations in imaginary time (whose integral gives ) imply strong correlations of the same kind for the current (while the opposite is not true, as current fluctuations may take place within the same topological sector).
It is tempting to associate the buildup of strong windingnumber correlations in imaginary time with the unbinding of VAV pairs along the same direction, leading to correlated windingnumber fluctuations. Fig. 9 shows indeed that close to the critical point the normalized winding number correlations
(21) 
appear to develop a tail that is consistent with a power law, unlike in the MI phase, and with an exponent significantly smaller than that which can be estimated in the SF phase. Our numerical results do not allow us to draw firm conclusions on the behavior of in the thermodynamic limit. Yet one may conjecture that, in the that limit and at , decays as a fast power law deep in the SF phase (in the absence of any gap in the spectrum); and that this powerlaw decay changes abruptly at the critical point, with the further appearance of an exponential decay related to the opening of the Mott gap. All these decay laws are expected to be integrable – given that the winding susceptibility keeps scaling linearly with across the transition, as imposed by the linear increase of – but their integral leads to strong corrections to scaling as in Eq. (19) when entering the MI phase.
The standard picture of the KT transition associates the unbinding of VAV pairs with the suppression of conventional correlations in real space – and the loss of coherence and stiffness as in Fig. 3 is certainly a striking consequence of this aspect. On the other hand, when focusing on windingnumber fluctuations, the unbinding of VAV actually leads to a critical enhancement of correlations for topologicalsector fluctuations, namely “topological correlations” are affected by topological defects in a rather different way from ordinary correlations. This point of view endows the jump in the superfluid stiffness in particular, and the 1 SF/MI transition in general, with a precise topological nature, related to the singular behavior of windingnumber correlations in imaginary time. The topological characterization of the SF/MI transition in the context of a 1 lattice Bose gas parallels the one recently provided for the thermal KT transition of the 2 lattice Coulomb gas in Ref. [15]; in the latter model, topological sectors are defined in terms of two winding numbers (one per dimension), each of which map to the structures in Fig. 1(a). The transition is accompanied by a total suppression of these classical topologicalsector fluctuations in the lowtemperature phase.
4 Thermal regime: Arrhenius onset of smooth phase slips
When turning on the temperature, windingnumber fluctuations can appear due to thermal excitations. Focusing our attention on the superfluid phase, we have seen in the previous section that phase configurations with largely dominate the ground state, and they are weakly admixed with configurations. The energy cost to create a finite for most of the configurations building up the ground state (or on most of the imaginarytime slices within the pathintegral representation in Sec. 2.2) can be easily estimated via the superfluid stiffness defined in Eq. (11). Indeed the least energetic configuration with corresponds to that minimizing the energy in the presence of twisted boundary conditions, which is achieved by distributing the phase twist uniformly along the ring, namely by developing a phase difference across each junction in the lattice. In the classical limit of the quantum phase model () the energy cost for such a configuration would be simply
(22) 
where and we consider the limit . Quantum fluctuations induced by , namely quantum phonons and QPS, renormalize this energy cost, which is properly estimated via the zerotemperature superfluid fraction
(23) 
Similarly the energy cost for a generic winding number can be estimated as , provided that . According to this estimate, one expects that thermal fluctuations produce configurations possessing uniformly distributed phase twists with an Arrhenius probability , leading to winding number fluctuations with variance
(24) 
where are multiplicity factors. Fig. 10(ac) shows interference patterns reconstructed from QMC snapshots bearing a finite winding number (up to ) in the regime of thermal excitation of lowenergy TPS on top of a superfluid ground state. Such lowenergy TPS result in very smooth spirals in the interference pattern, decorated with localized dislocations of the fringe pattern due to quantum fluctuations. Moreover, for sufficiently large one often observes configurations exhibiting both a TPS and a QPS, clearly distinguished by the extended nature of the former and the localized nature of the latter (see Fig. 10(d) for an example, as Sec. 5 for further discussion).
The Arrhenius law for winding number fluctuations in Eq. (24) immediately implies an exponential sensitivity to the system size, and the general scaling behavior
(25) 
which is obviously inconsistent with the scaling behavior of Eq. (14) valid in the quantum regime. As shown in Fig. 11 such a scaling form is indeed very well verified by our data down to a characteristic crossover temperature at which the thermal regime breaks down, leaving space to the quantum regime which rather obeys the Poissonian scaling as in Eq. (14). The detailed study of the crossover will be the subject of Sec. 5.
Moreover Eq. (24) implies an exponential temperature dependence of the thermal windingnumber fluctuations, dominated by the term. This exponential dependence is indeed clearly observed in our data within the SF regime (), as shown in Fig. 12(a): in particular when the exponential decrease of dominates the intermediate temperature range , but it then levels off at a  (and size) dependent temperature, marking the crossover to the regime of quantum windingnumber fluctuations, insensitive to temperature. As predicted by Eq. (24) the exponential decay rate of is controlled by the zerotemperature superfluid density, as it is very well verified by the data in Fig. 12(b). This in turn shows that the study of the thermal dependence of the windingnumber fluctuations is an exponentially sensitive probe of the superfluid density; we will come back to this point in the Conclusions.
So far we have identified the thermal regime of windingnumber fluctuations with: 1) the peculiar scaling form it satisfied by fluctuations as in Eq. (25); 2) the exponential activation of fluctuations controlled by the (zerotemperature) superfluid fraction; and 3) by the uniformly twisted phase configurations that are thermally populated in the finite sector at sufficiently low temperatures. This latter aspect reflects itself not only in the qualitative aspects of the experimentally accessible interference patterns (as shown in Fig. 10), but also (and more quantitatively) in a peculiar structure for the distribution of current fluctuations, which are accessible experimentally via the interferometric reconstruction of the phase differences across the tunnel barriers. Indeed a smooth phase twist with winding number displays phase differences among neighboring sites, leading to a current . If one limits oneself to smooth phase twists only, current fluctuations should develop a discrete distribution with peaks corresponding to the discrete values. Such a picture is indeed realized in the regime of weak interactions and low temperatures, where nonsmooth twists are prevented from forming both quantummechanically and thermally – see Fig. 13; on the other hand, upon increasing the multipeak structure is rounded off when the occurrence of localized QPS equals and eventually overcomes that of the smooth TPS, leading to unconstrained current fluctuations which progressively develop into a Gaussian distribution.
5 Quantumtothermal crossover and thermal revival of quantum phase slips
5.1 Quantumtothermal crossover, and thermal revival of quantum fluctuations, from the quantumtototal fluctuation ratio
As pointed out in Secs. 3 and 4, thermal and quantum windingnumber fluctuations display a rather different scaling behavior, and they possess as well rather different spatial properties. Hence a most interesting aspect is to locate the crossover between the two regimes, and to determine its most important signatures. To achieve the first goal unambiguously, we can rely again on the pathintegral formulation of the partition function for the quantum phase model discussed in Sec. 2.2, identifying quantumcoherent windingnumber fluctuations with those taking place along the imaginarytime dimension. In contrast, thermal (or incoherent) fluctuations do not involve the imaginarytime direction, but they originate from fluctuations between different pathintegral configurations contributing to the partition function. In order to eliminate the information on the imaginarytime fluctuations one can define the imaginarytimeaveraged winding number
(26) 
and hence define the thermal windingnumber fluctuations as those of :
(27) 
It is easy to prove that , namely the total fluctuations always exceed the thermal ones, because of the very existence of imaginarytime fluctuations. Hence quantum winding number fluctuations can be estimated as the residual fluctuations [18]:
(28) 
In the ground state, namely in the limit in which the imaginarytime direction becomes infinitely long, each pathintegral configuration should sample the whole statistics of fluctuations associated with the quantum ground state ; this implies that due to timereversal symmetry. As a consequence the configurationtoconfiguration fluctuations vanish, and, as expected, quantum fluctuations as in Eq. (28) saturate the total fluctuations. The fluctuations of are instead thermally activated; as a consequence, we can evaluate the relative strength of quantum fluctuations with the ratio
(29) 
satisfying the two limiting behaviors for and for . It is then natural to define a (size and parameterdependent) crossover temperature between the quantum and thermal regime of winding number fluctuations as the one satisfying the criterion:
(30) 
Fig. 14 shows as a function of for different values of the interaction strength: the crossover temperature systematically moves to higher values as increases, as expected due to the increase of the strength of quantum windingnumber fluctuations.
Beside the growth of the crossover temperature with interaction strength, one can observe that in the SF phase a highly nontrivial interplay between thermal and quantum windingnumber fluctuations is exhibited by the data when looking at a broad temperature scale. Indeed is found to display a nonmonotonic behavior in the SF phase, with a further maximum at finite , shifting to lower temperatures as increases. This feature is found not to depend on the value of we chose (namely it is a feature of the quantum phase model as such), and, within the SF phase, it is also not dependent on the size, as shown in Fig. 18. A possible interpretation of this peculiar thermal revival of quantum fluctuations stems from the mapping of the 1 quantum phase model onto the anisotropic 2 XY model of Eq. (4), suggesting that quantum windingnumber fluctuations are induced by the appearance of VAV excitations separating along the imaginarytime dimension (Fig. 1(b)). As we have seen in Sec. 2.1, at fixed all couplings constants are held fixed and the temperature uniquely controls the length of the extra dimension, . The thermal shrinking of drives the system from fully 2 at to quasi1 at finite temperature, and this dimensional crossover alone is found to be sufficient to enhance the vorticity in the system. We can quantitatively relate the thermal enhancement of quantum windingnumber fluctuations to the proliferation of VAV pairs by calculating the squared vortex density of the fictitious 2XY model of Eq. (4)
(31) 
where is the function defined in Eq. (7), runs over the spacetime square plaquettes and runs counterclockwise over the four sites of each plaquette. This quantity is an exclusive property of the pathintegral representation in discrete imaginary time, and it does not correspond to a physical observable of the quantumphase model (in fact, it is found to depend on the choice of ). Yet, as shown in Fig. 15, its temperature derivative exhibits a maximum which correlates very well with the finite maximum of the , slightly shifting to lower temperatures when the interaction increases. The temperature dependence of stems entirely from the dimensionality crossover described above: as the system moves from a torus geometry to an annulus geometry upon decreasing M, its stiffness (in the sense of the XY model of Eq. (5)) opposing to the thermal activation of VAV pairs is gradually reduced – given that, all other parameters being held fixed, the stiffness of an XY model is higher in than in .
5.2 Crossover from measurable quantities
The quantumtothermal crossover in winding number fluctuations at low temperature, and the unusual phenomenon of thermal enhancement of quantum fluctuations at intermediate temperature, is not only revealed theoretically by the ratio, but it is also witnessed directly by measurable quantities. In the Appendix (Sec. 8) we show representative interference patterns obtained from uncorrelated snapshots of the MC simulation, mimicking the outcomes of uncorrelated shots of the experiment. In contrast to that found at (see Fig. 4) we observe that increasing the temperature above zero first leads to the appearance of uniform phase slips leading to smooth spirals in the interference patterns, corresponding to TPS which fully dominate the pathintegral configuration at all imaginarytime slices; only rarely does one observe localized phase slips (both in space and in imaginary time), corresponding to quantum fluctuations – this is consistent with the initial drop of . On the other hand, as the temperature grows further the localized phase slips are seen to be increasingly frequent, in accordance with the thermal increase of .
We can make this qualitative observation on the geometry of the interference fringes much more quantitative by defining the distance of finite phase configurations at fixed imaginary time, , from a perfectly uniform phase slip with the same , namely
(32) 
where represents the thermal average restricted to the topological sector with winding number . The above quantity looks for the minimal distance between the phase configuration and a configuration exhibiting a uniform phase twist of , starting from any site on the lattice (here is defined modulo ). In the MC simulation we then average the distance over all configurations exhibiting the same finite winding number (in absolute value).
Another sensitive observable capturing the quantumtothermal crossover is given by the integrated current correlations at finite windingnumber
(33) 
restricted to the topological sector with winding number . A uniform phase slip with winding number exhibits strong currentcurrent correlations, ; on the other hand, such correlations are suppressed if the phase slip is localized: in the limit of a phase slip localized e.g. on three bonds (with a slip of each), . For and , in the first case, while in the second case.
The average and are plotted in Fig. 16, showing that the quantumtothermal crossover at is marked by an abrupt drop of the distance to a uniform phase slip, accompanied by a sudden enhancement of currentcurrent correlations. This shows that there is a fundamental correspondence between the coherent nature of the phase slips leading to windingnumber fluctuations, and the spatial structure of such excitations, which is fully accessible to experiments. Moreover, the two quantities under investigation show a broad minimum and a plateau (respectively) over the temperature range in which nearly vanishes, marking the thermal regime. On the other hand, upon increasing the temperature shows a revival, and a suppression, both corresponding to the revival in the . The enhancement of quantum windingnumber fluctuations is certainly responsible for the temperature dependence of and , but one should add to this the contributions of thermally populated phonons which, while not affecting at all , are certainly responsible for a distortion of smooth spirals with respect to the low regime, and for a weakening of currentcurrent correlations.
5.3 Interaction and sizedependence of the crossover temperature; phase diagram
We conclude the discussion of the quantumtothermal crossover by focusing on its dependence on the system size and interaction strength. In the SF case such dependence can be readily extracted by identifying the crossover temperature with the one at which thermal windingnumber fluctuations, setting in with the Arrhenius law of (24), overcome the quantum fluctuations, namely
(34) 
which, when solving for , gives
(35) 
Hence we observe that the crossover temperature vanishes as (up to logarithmic corrections – see also Fig. 18(a)), and that its interaction dependence stems from a competition between the interaction dependence of and . Indeed, while tends to vanish upon increasing , the groundstate fluctuations increase above zero, eliminating the divergence of in the denominator. It turns out that the latter effect wins for weak interactions, given also the exponential onset of the quantum windingnumber fluctuations following Eq. (17) (see also Fig. 6(b)).
Fig. 17 shows that, sufficiently deep in the SF phase, Eq. (34) provides an excellent estimate of the finitesize crossover temperature when compared with the one satisfying the criterion of Eq. (30). Eq. (35) clearly shows that the quantumtothermal crossover in the SF regime is a mesoscopic effect, and that the crossover temperature is pushed to zero as due to the collapse of the finite topological sectors onto the sector which dominates the ground state of finitesize systems. The size dependence of the crossover temperature, along with the size dependence of quantum windingnumber fluctuations (Eq. (14)) points towards the existence of an optimal size range for the experimental observation: indeed, while a small size enhances the crossover temperature, the number of measurements required to observe Poissonian quantum windingnumber fluctuations decreases with .
On the contrary, in the MI regime the crossover temperature remains finite even in the thermodynamic limit, as shown in Fig. (18)(b). The ground state already features a strong coherent admixture of different topological sectors, and ground state physics is protected by the Mott gap: windingnumber fluctuations therefore maintain quantum coherence up to a temperature which should be comparable with the gap itself. In Fig. 17 we sketch therefore the crossover temperature in the thermodynamic limit, vanishing at the SF/MI quantum critical point, and building up as in a manner similar to the Mott gap.
6 Conclusions
In this paper we have shown that winding number fluctuations in lattice Bose rings display a very rich behavior, exhibiting a quantum regime of localized quantum phase slips, and a thermal regime of extended thermal phase slips. The fundamental correspondence between the thermal or quantum nature of phase slips, and their spatial structure, makes this crossover detectable in stateoftheart experiments on ultracold gases, which are able to fully reconstruct the phase configuration of the ringtrapped Bose gas via matterwave interference. It is also worth noticing that the windingnumber fluctuations stand as a direct probe of the groundstate superfluid density, even without probing the response of the system to rotation or to a gauge field. Indeed, by looking at the thermal windingnumber fluctuations one restricts naturally to topologically nontrivial sectors of the phase configurations, in which the system has effectively come into rotation spontaneously. The groundstate superfluid density is exactly what opposes the appearance of such fluctuations, and its renormalization due to quantum fluctuations is directly observable as a softening of thermal windingnumber fluctuations.
Our work defines the experimental conditions for the direct observation of quantum phase slips in atomcircuit setups, such as the ones explored in Refs. [10, 12]. This perspective nicely complements the recent effort in the context of superconducting nanostructures [27, 25, 5, 14], fully exploiting the power of atom interferometry. Further extensions of our work can be readily envisioned to involve topological excitations in higherdimensional systems, or the study of quantumquench dynamics across the superfluidinsulator transition.
7 Ackowledgements
Fruitful discussions with A. Rançon are gratefully acknowledged. This work is supported by ANR (‘ArtiQ’ project).
8 APPENDIX: Interference patterns at the quantumtothermal crossover
In this appendix we present representative interference patterns with , stemming from uncorrelated snapshots of our Monte Carlo simulations on a ring with weak interactions . Such patterns mimick the outcome of independent experimental runs obtained under the same conditions. Figs. 19, 20 and 21 show interference patterns for a growing temperature, and they establish a link between the appearance of localized dislocations in the interference fringes, and their origin from quantumcoherent fluctuations of the winding number.
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