Abstract
Topology in quantum manybody systems has profoundly changed our understanding of quantum phases of matter. The model that has played an instrumental role in elucidating these effects is the antiferromagnetic spin1 Haldane chain^{1,2}. Its ground state is a disordered state, with symmetryprotected fourfolddegenerate edge states due to fractional spin excitations. In the bulk, it is characterized by vanishing twopoint spin correlations, gapped excitations and a characteristic nonlocal order parameter^{3,4}. More recently it has been understood that the Haldane chain forms a specific example of a more general classification scheme of symmetryprotected topological phases of matter, which is based on ideas connected to quantum information and entanglement^{5,6,7}. Here, we realize a finitetemperature version of such a topological Haldane phase with Fermi–Hubbard ladders in an ultracoldatom quantum simulator. We directly reveal both edge and bulk properties of the system through the use of singlesite and particleresolved measurements, as well as nonlocal correlation functions. Continuously changing the Hubbard interaction strength of the system enables us to investigate the robustness of the phase to charge (density) fluctuations far from the regime of the Heisenberg model, using a novel correlator.
Main
Topological phases of matter often share a deep connection between their bulk and edge properties^{8,9}. In the case of the Haldane chain, the bulk exhibits a hidden antiferromagnetic (AFM) order characterized by AFM correlations interlaced with an arbitrary number of S^{z} = 0 elements, where S^{z} denotes the spin component in the zdirection. This pattern can only be revealed through nonlocal string correlations that are sensitive to the local spin states, which, however, require detection of the quantum manybody system with microscopic resolution. Although this was not possible in early experiments on spin1 chains, evidence for a spin gap, as well as spin1/2 localized edge states, was found using neutron scattering^{10,11} or electron resonance experiments^{12,13} while not directly probing this hidden order or spatially resolving the edge states. Recent developments in quantum simulations enable one to go beyond such solidstate bulk measurements by observing quantum manybody systems with singlesite resolution^{14,15,16,17,18} and in a fully spin and densityresolved way^{19,20}. This provides a rich diagnostic tool to obtain a direct microscopic picture of the hidden order in experiments^{21,22}. The power of this technique has also been demonstrated recently in a study that was able to reveal a symmetryprotected topological (SPT) phase in the hardcore boson Su–Schrieffer–Heeger (SSH) model using Rydberg atoms^{23}. Here we expand on those results by realizing a finitetemperature version of the Haldane phase in a spin system with tuneable coupling strength, system size and controlled charge fluctuations. We show this by measuring both topological and trivial string order parameters.
An instructive way to engineer the Haldane phase in systems of spin1/2 fermions is on the basis of the AKLT model^{4,24}, in which a spin1 particle is formed out of two spin1/2 particles. Thus, spin1/2 ladder systems emerge as an experimentally realizable platform for the Haldane phase. Whereas a natural implementation with spin1 particles on individual rungs requires ferromagnetic rung couplings and antiferromagnetic leg couplings, a much wider variety of couplings in spin1/2 quantum ladders feature the Haldane phase^{25,26}. This includes the antiferromagnetic Heisenberg case, which we realize here as the stronginteraction limit of the Fermi–Hubbard model.
In our experiment, we prepare such ladders by adiabatically loading a spinbalanced mixture of the two lowest hyperfine states of ^{6}Li into an engineered lattice potential (Methods). As illustrated in Fig. 1a, we realize four isolated twoleg ladders with a variable number (L) of unit cells (where L is therefore also equivalent to length), surrounded by a lowdensity bath of particles^{27}. The unit cells are chosen to be either along the rungs of the ladders (vertical unit cell, Fig. 1b) or along the diagonals (diagonal unit cells, Fig. 1c). The edges of the ladders are then engineered to match the choice of unit cell: straight edges are chosen for vertical unit cells, whereas one site is blocked on each edge in the case of diagonal unit cells. The atoms in the lowest band of the optical lattice are well described by the Fermi–Hubbard model, with tunnelling energies, \({t}_{\parallel }\) (chain), \({t}_{\perp }\) (rung), and onsite interactions U. For halffilling and at strong \(U/{t}_{\parallel ,\perp }\approx 13\), used throughout most of our experiments (see Methods for details), density fluctuations are suppressed and the spin ladder realizes the Heisenberg model^{28} with Hamiltonian:
with positive leg and rung couplings, \({J}_{\parallel ,\perp }=4{t}_{\parallel ,\perp }^{2}/U\) and the spin1/2 operators \({\hat{{\bf{S}}}}_{x,y}\) at site (x, y), with A, B denoting the two legs of the ladder.
The topological properties are most easily explained in the limit \({J}_{\perp }\gg {J}_{\parallel }\), where strong spin singlets form along the rungs and the system exhibits an energy gap of \({J}_{\perp }\). The behaviour on the edges of the ladder then depends on how the system is terminated, and therefore on which unit cells have been chosen. For diagonal unit cells (Fig. 1b), two unpaired spin1/2 particles remain and the manybody system has a fourfold degeneracy that is only weakly lifted by an edgetoedge coupling, which vanishes exponentially with system size (Supplementary Information). In the trivial case of vertical unit cells (Fig. 1c), all spins pair into singlets and the ground state is unique. These descriptions remain valid even for weaker \({J}_{\perp }/{J}_{\parallel }\), where the singlet alignment may change between vertical and horizontal, but any line between two rungs cuts an even number of singlets^{29,30}.
To make the analogy between the spin1/2 ladder and the Haldane integer chain more apparent, we switch to a description in terms of total spin per kth unit cell, \({\hat{{\bf{S}}}}_{k}={\hat{{\bf{S}}}}_{k,{\rm{A}}}+{\hat{{\bf{S}}}}_{k,{\rm{B}}}\), where the indices (A, B) indicate the two spin1/2 particles in the same unit cell, making an integer spin. In the diagonal unit cell such a system shows a high (≥80%) triplet fraction^{26} (Supplementary Information). We note that this spin ladder can be adiabatically connected to a spin1 chain by including ferromagnetic couplings within the unit cell^{25}. However, having a high triplet fraction is not essential for having a welldefined Haldane phase, as both systems share the same universal SPT features^{26}.
The defining property of the Haldane SPT phase is that it is an integerspin chain with spin1/2 edge modes: the bulk SO(3) symmetry is said to fractionalize into SU(2) symmetry at the edge. It has no spontaneous symmetry breaking and thus the spin correlation function \(\langle {\hat{S}}_{k}^{z}{\hat{S}}_{k+d}^{z}\rangle \) is short range. Instead, the aforementioned symmetry fractionalization^{6,7} can be detected in the bulk using string order parameters^{3,31}:
with an onsite symmetry \({\hat{U}}_{l}\) and endpoint operator \({\hat{{\mathscr{O}}}}_{k}\), where l denotes the unit cell and d the string distance (Fig. 2 and Supplementary Information). This correlator probes the transformation behaviour of the bulk under a symmetry \({\hat{U}}_{l}\); for example, a spin rotation around the z axis by π, \({\hat{R}}_{l}^{z}\equiv {\rm{\exp }}({i}{\rm{\pi }}{\hat{S}}_{l}^{z})\). The purestring correlator \({g}_{{\mathbb{1}},{R}^{z}}(d)\), where \({\hat{{\mathscr{O}}}}_{k}=1\) and \({\hat{U}}_{l}={\hat{R}}_{l}^{z}\), is nonzero for \(d\gg 1\) if the edge does not have halfinteger spins^{31}. This is the case for the topologically trivial configuration but not for the Haldane phase, in which the symmetry is fractionalized. The spinstring operator \({g}_{{S}^{z},{R}^{z}}(d)\)(ref. ^{3}), \({\hat{{\mathscr{O}}}}_{k}={\hat{S}}_{k}^{z}\), exhibits the opposite behaviour and is nonzero only in the Haldane phase (see Supplementary Information for details about the symmetries of the Haldane phase). Thus we can identify the Haldane phase by comparing the two string correlators \({g}_{{S}^{z},{R}^{z}}\) and \({g}_{{\mathbb{1}},{R}^{z}}\), and observe opposite behaviour in the two different regimes.
Another perspective on \({g}_{{S}^{z},{R}^{z}}\) can be gained by recognizing it as a normal twopoint correlator at distance d, which ignores all spin0 contributions along the way (‘squeezed space’^{22,32}). In the underlying spin1/2 system, this order stems from N − 1 consecutive rungs dominantly consisting of N − 1 singlets and two spin1/2 states, which have a combined total spin of +1, 0 or −1.
To observe the characteristics of the SPT phase, we prepare a twoleg ladder of length L = 7 and \({J}_{\perp }/{J}_{\parallel }=1.3(2)\) in both the topological and the trivial configuration. The tailored potential yields a homogeneous filling of the system with sharp boundaries (Fig. 2a), which is characterized by a remaining density variance over the system of 2 × 10^{−4}. To focus on the spin physics, we select realizations with \({N}_{\uparrow }+{N}_{\downarrow }=2L\) per ladder. Additionally, we exclude ladders with an excessive number of doublon–hole fluctuations and do not consider strings with odd atom numbers in the string or the endpoints of the correlator (Methods). We characterize the spinbalanced ladder systems \(({M}^{z}\equiv ({N}_{\uparrow }{N}_{\downarrow })/2=0)\) by evaluation of the string order parameters, as defined in equation (2). In the topological configuration, we observe fast decay of \({g}_{{S}^{z},{R}^{z}}\) over a distance of approximately one site and a longrange correlation up to d = 6, with a final value of \({g}_{{S}^{z},{R}^{z}}\simeq 0.1\) (Fig. 2b). In contrast, for the trivial configuration, the correlation decays rapidly to zero as a function of the string correlator length. The opposite behaviour is seen for \({g}_{{\mathbb{1}},{R}^{z}}(d)\), demonstrating the hidden correlations expected for both phases.
Furthermore, the twopoint spin correlation, \(C(d)\equiv {g}_{{S}^{z},{\mathbb{1}}}(d)=\)\(\langle {\hat{S}}_{k}^{z}{\hat{S}}_{k+d}^{z}\rangle \,,\) yields only the shortrange AFM correlation characteristic for a gapped phase (see insets in Fig. 2b). It is interesting to note that at the largest distance in the topological case, C(d = 6) displays a clear (negative) correlation between the two edge spins, despite small correlations at shorter distances. This (classical) correlation confirms the existence of a nonmagnetized bulk, such that spins on the edges of the system must be of opposite direction at global M^{z} = 0.
We probe the edges explicitly by measuring the amplitude of the local rungaveraged magnetization m^{z}(x) as a function of rung position x for different sectors of the ladder magnetization M^{z} (Fig. 2c). In the case of an imbalanced spin mixture with M^{z} = ±1, we see that the two end sites exhibit on average a higher magnetization than the bulk rungs in the topological configuration. This is consistent with the bulk of the ground states of both phases forming a global singlet, and only the edges of the topological phase carrying an excess spin1/2 without energy cost. The measured bulk magnetization can be attributed to finitetemperature effects (Supplementary Information).
The SPT phase is expected to be robust^{26} on variation of the ratio \({J}_{\perp }/{J}_{\parallel }\), but maintains a finite gap in the system. We realize both the trivial and topological configuration with different \({t}_{\perp }/{t}_{\parallel }\) at almost fixed U and study the string correlators at maximal distance (L − 1) versus \({J}_{\perp }/{J}_{\parallel }\) (Fig. 3a). For the topological configuration with diagonal unit cells, we observe \({g}_{{\mathbb{1}},{R}^{z}}(L1)\simeq 0\) and \({g}_{{S}^{z},{R}^{z}} > 0\) for all \({J}_{\perp }/{J}_{\parallel }\) with a maximum around \({J}_{\perp }/{J}_{\parallel }\simeq 1.3(2)\) (that is, \({J}_{\perp }/({J}_{\perp }+{J}_{\parallel })\simeq 0.56(4)\)), whereas for the trivial case the role of the correlators is reversed. Both phases continuously connect in the limit of two disconnected chains at \({J}_{\perp }=0\). These observations demonstrate qualitatively all the key predictions of the antiferromagnetic spin1/2 ladder at temperature T = 0 (ref. ^{26}) and the strengths of the measured correlations are consistent with exact diagonalization (ED) calculations using an entropy per particle \(S/N=(0.30.45)\,{k}_{{\rm{B}}}\) (shaded lines in Fig. 3a).
We reveal these features despite the finite temperature in our system, which would destroy the longrange hidden order in an infinite system. The total entropy in our system is, however, still low enough to yield a large fraction of realizations of the topological ground state. In larger systems, the total number of thermal excitations grows (at fixed entropy per particle) and the nonlocal correlator \(\,{g}_{{S}^{z},{R}^{z}}(L1)\) decreases (see inset of Fig. 3a), consistent with vanishing correlations in the thermodynamic limit, thus yielding a restriction on our system size at our level of experimental precision and entropy per particle (Supplementary Information). Finite size effects are explored in detail in the Supplementary Information. We note that the difference between the SPT phase and the trivial phase is here clearly shown by considering both \({g}_{{S}^{z},{R}^{z}}\) and \({g}_{{\mathbb{1}},{R}^{z}}\).
To investigate the localization length of the edge states, we evaluate our data for \({m}^{z}\) = ±1 and plot the local magnetization per unit cell \({m}^{z}\)(k) for different system sizes (Fig. 3b). Because of the singlets in the bulk, the excess spin is most likely to be found at the edges of the system. This spin partly polarizes the neighbouring sites antiferromagnetically, leading to an exponentially localized net magnetization with AFM substructure^{33}. The data are well described by the fit function \({m}^{z}(k)={m}_{{\rm{B}}}+{m}_{{\rm{E}}}\left({(1)}^{k}{{\rm{e}}}^{k/\xi }+{(1)}^{Lk1}{{\rm{e}}}^{(Lk1)/\xi }\right)\) with free bulk magnetization m_{B}, edge magnetization m_{E} and decay length ξ. In Fig. 3c, we show how this localization length ξ decreases as we approach the limit of rung singlets, \({J}_{\perp }\gg {J}_{\parallel }\). Comparison with ED lets us identify two regimes: at \({J}_{\perp }\gtrsim {J}_{\parallel }\), the measured ξ drops with larger \({J}_{\perp }\) and coincides with theory independent of temperature, whereas at low \({J}_{\parallel }\) thermal effects dominate, limiting the increase of ξ to three sites for our system (Supplementary Information).
Thus far, we have worked in the Mott limit where density fluctuations can be ignored, such that the spin Hamiltonian, equation (1), is a good effective description of the Fermi–Hubbard ladder. However, it is known that the Haldane SPT phase can be unstable to density fluctuations^{34,35,36}. By reducing \(U/{t}_{\parallel }\), the symmetry in the unit cell in the bulk changes from SO(3) to SU(2), as it now may contain both halfinteger and integer total spin. This effectively removes the distinction between bulk and edge (Supplementary Information). This means that the edge mode and string order parameter are no longer topologically nontrivial, which is also manifested in the fact that the two phases can be adiabatically connected by tuning through a low\(U/{t}_{\parallel }\) regime if one breaks additional symmetries but preserves spinrotation symmetry^{34,35,36}. In particular, the above string orders lose their distinguishing power: \({g}_{{S}^{z},{R}^{z}}\) and \({g}_{{\mathbb{1}},{R}^{z}}\) will both generically have longrange order away from the Mott limit^{34}.
Intriguingly, despite the breakdown of the above symmetry argument and string order parameter, the Hubbard ladder (with diagonal unit cell) remains a nontrivial SPT phase due to its sublattice symmetry. This symmetry is a direct consequence of the ladder being bipartite (see Supplementary Information for details). It is simplest to see that this protects the SPT phase in the limit U = 0, where the two spin species decouple, such that our model reduces to two copies of the SSH chain^{37}. It is known that such a stack remains in a nontrivial SPT phase in the presence of interactions, namely U ≠ 0 (ref. ^{38}). Moreover, together with the parity symmetry of spindown particles, \({\hat{P}}_{l}^{\downarrow }\equiv {\rm{\exp }}\left[{i}{\rm{\pi }}\left({\hat{n}}_{l,{\rm{A}}}^{\downarrow }+{\hat{n}}_{l,{\rm{B}}}^{\downarrow }\right)\right]\), it then gives rise to a different string order parameter: the topological phase is characterized by longrange order in \({g}_{{S}^{z},{P}^{\downarrow }}\), whereas it has vanishing correlations for \({g}_{{\mathbb{1}},{P}^{\downarrow }}\), with the roles being reversed in the trivial phase. This novel string order is derived in the Supplementary Information. Remarkably, in the Heisenberg limit, it coincides with the conventional string order parameter used before.
In the topological phase it is meaningful to normalize \({g}_{{S}^{z},{P}^{\downarrow }}\) to \({\tilde{g}}_{{S}^{z},{P}^{\downarrow }}=\eta {g}_{{S}^{z},{P}^{\downarrow }}\) with \({\eta }^{1}=\langle {\hat{S}}_{k}^{z}{\hat{S}}_{k+d}^{z}\rangle \), which effectively excludes endpoints with spin S^{z} = 0. Indeed, we find unchanged string correlations \({\tilde{g}}_{{S}^{z},{P}^{\downarrow }}\) and \({g}_{{\mathbb{1}},{P}^{\downarrow }}\) down to the lowest experimentally explored value \(U/{t}_{\parallel }=2.5(2)\) (Fig. 4a, b) and edge state signals down to \(U/{t}_{\parallel }=5.0(3)\) (Fig. 4c). Density matrix renormalization group (DMRG) calculations for \(L\to \infty \) confirm nonzero \({\tilde{g}}_{{S}^{z},{P}^{\downarrow }}\) (L−1) at T = 0 and for all rungcoupling strengths (Fig. 4d), while \({g}_{{\mathbb{1}},{P}^{\downarrow }}(L1)\) is strictly zero. Owing to the normalization \({\tilde{g}}_{{S}^{z},{P}^{\downarrow }}(L1)\) goes to 1 for \({J}_{\perp }\gg {J}_{\parallel }\).
In our work, we realized a finitetemperature version of the topological Haldane SPT phase using the full spin and density resolution of our Fermi quantum gas microscope. We demonstrated the robustness of the edge states and the hidden order of this SPT phase in both the Heisenberg and the Fermi–Hubbard regime. In the future, studies may extend the twoleg ladder to a varying number of legs, in which one would expect clear differences between even and odd numbers of legs^{39} and topological effects away from halffilling^{40}, or may investigate topological phases in higher dimensions^{41}. Furthermore, the ladder geometry holds the potential to reveal hole–hole pairing^{42} at temperatures more favourable than in a full twodimensional system.
Methods
Experimental sequence
In each experimental run, we prepare a cold atomic cloud of ^{6}Li in a balanced mixture of the lowest two hyperfine states (F = 1/2, m_{F} = ±1/2). For evaporation, we confine the cloud in a single layer of a staggered optical superlattice along the z direction, with spacings a_{s} = 3 μm and a_{1} = 6 μm and depths \({V}_{s}=45\,{E}_{{\rm{R}}}^{{\rm{s}}}\) and \({V}_{l}=110\,{E}_{{\rm{R}}}^{{\rm{l}}}\), where E_{R} denotes the recoil energy of the respective lattice. The atoms are harmonically confined in the xy plane and the evaporation is performed by ramping up a magnetic gradient along the y direction^{20}. The final atom number is tuned via the evaporation parameters.
The cloud is then loaded into an optical lattice in the xy plane with spacings a_{x} = 1.18 μm and a_{y} = 1.15 μm, which is ramped up within 100 ms to its final value, ranging from 5E_{R} to 15E_{R} depending on the chosen Hubbard parameters. The scattering length is tuned from 230 a_{B} during evaporation, with a_{B} being the Bohr radius, to its final value ranging between 241 a_{B} and 1,200 a_{B} using the broad Feshbach resonance of ^{6}Li. An overview of the parameters of each dataset is given in Extended Data Table 1. Simultaneously with the lattice loading, a repulsive potential is ramped up, which compensates for the harmonic confinement generated by the curvature of the Gaussian lattice beams and divides the resulting flat area into four disconnected ladder systems along the y direction (see ‘Potential shaping’). We achieve temperatures of \({k}_{{\rm{B}}}T\approx 0.9(3){J}_{\parallel }\) for the parameters in Fig. 2.
For detection, the configuration is frozen by ramping the xy lattices to \(43\,{E}_{{\rm{R}}}^{xy}\) within 250 μs. A Stern–Gerlach sequence separates the two spin species into two neighbouring planes of the vertical superlattice, which are then separated to a distance of 21 μm using the charge pumping technique described in ref. ^{20}. Finally, simultaneous fluorescence images of the two planes are taken using Raman sideband cooling in our dedicated pinning lattice^{43}, with an imaging time of 2.5 s. The fluorescence of both planes is collected through the same highresolution objective. The light is then split into two paths using a polarizing beam splitter. One of the beams passes through a variable 1:1 telescope before both paths are recombined on a second polarizing beam splitter with a small spatial offset. This enables us to image both planes in a single exposure, with each plane in focus on a separate fixed position of our camera. We calibrated the relative position on the camera of the two imaged planes using 300 shots of a spinsplit Mott insulator and the matching algorithm described in the supplement of ref. ^{20}. The overall detection fidelity per atom was 96(1)%.
Potential shaping
The ladder systems are created by superimposing the optical lattice with a repulsive potential, which is shaped by projecting incoherent light at 650 nm (generated from a SLED by Exalos EXS21003003) from a digital micromirror device (DMD) through the highresolution objective. Four ladders are created by blocking lattice sites with a potential V = 3.5(5)U around each ladder. The area outside the walled ladders is lifted above the inner ladder potential, but remains below the interaction energy U. It thus serves as a reservoir for surplus atoms, which occupy this region once the lowest Hubbard band of the ladders is filled. The flatness of the potential is adjusted for each parameter setting, as the intensity of the lattice beams directly influence the curvature of the potential. This is accomplished by realizing a system with about 20% doping and returning the average density of 100–150 experimental runs as feedback to the DMD pattern. We repeat the feedback until we reach a sufficiently flat density profile with a variance <1 × 10^{−3} over the 8L ladder sites. To adjust for drifts in the lattice phase, we continuously track the lattice phase of each experimental run and feedback to the potential position accordingly. In Extended Data Fig. 1, the average density and occupation histograms of all four ladders and the reservoir area are shown for the dataset of L = 7.
Data selection
In each experimental run, four ladder systems are realized. To fulfil the criteria of the Heisenberg regime, we then select on ladder instances with atom number N = 2L and restrict the total magnetization to \({M}^{z}\) = 0, \({M}^{z}=1\), or \({M}^{z}\le 1\), depending on the observable, and specify the magnetization sector whenever data points are presented. \({M}^{z}\le 1\) for 87.5% of all data. Additionally, for all measurements in the Heisenberg regime, we remove ladders with more than two doublons, as those indicate a mismatch of the DMD pattern relative to the lattice phase. To give a specific example, we here give the precise numbers for the data presented in Fig. 2. This dataset consists of 7,032 realizations with four ladders each. Out of those 28,128 ladders, 6,721 have an atom number of 14. In addition, 2,636 ladders then have a total magnetization \({m}^{z}\) = 0 and 3,094 have a magnetization of \({M}^{z}\) = ± 1. Finally 77 of those 2,636 ladders have more than two doublon–hole pairs, which we exclude as these are most likely to be caused by drifts of the potential pattern given by the DMD. This leaves a total of 2,559 ladders out of the initial 28,128 for calculation of the string correlator.
For calculating the string correlators \({g}_{{S}^{z},{R}^{z}}\) and \({g}_{{\mathbb{1}},{R}^{z}}\) at fixed d, we exclude realizations with an odd atom number in the bulk area (grey area in the cartoon of Fig. 2b), as those would lead to imaginary contributions to the correlators, and we also exclude odd atom numbers at the edge areas (green in the cartoon of Fig. 2b). These cases are mostly due to the finite \(U/{t}_{\parallel }\), which still allows for some particle fluctuations. We keep other particle–hole fluctuations, such as those occurring within the string. These do not alter the observed string correlation relative to the Heisenberg model.
Nearestneighbour spin correlations
The twoleg ladder systems show strong antiferromagnetic spin correlations in which the dominant orientation depends on the ratio of couplings \({J}_{\perp }/{J}_{\parallel }\) and the strength is measured by \({C}_{x}(d)=4\langle {\hat{S}}_{i,j}^{z}{\hat{S}}_{i,j+d}^{z}\rangle \) and \({C}_{y}=4\left\langle {\hat{S}}_{{\rm{A}},j}^{z}{\hat{S}}_{{\rm{B}},j}^{z}\right\rangle \). For a leg coupling \({J}_{\parallel }\) much higher than the rung coupling \({J}_{\perp }\), the nearestneighbour spin correlator C_{y} along the rung almost vanishes, whereas correlations reach C_{x}(1) = −0.500(6) along the leg direction. For a dominating rung coupling \({J}_{\perp }\), C_{y} reaches −0.58(1), indicating singlet formation between the two sites of a rung. Extended Data Fig. 2a shows the measured spin correlations along both rung and leg for different values of \({J}_{\perp }/{J}_{\parallel }\). The values match the finitetemperature Heisenberg model for an entropy of S/N = (0.3−0.4) k_{B} per particle obtained from ED.
Theory simulation
In this work, we have used two different numerical methods to obtain theoretical predictions for the experimentally measured observables. The results in the Heisenberg regime were obtained using ED of our spin1/2 ladders up to sizes of L = 9 (limited by computational resources). For each data point, the system size and geometry in the ED simulation are the same as in the experimental data. The finitetemperature results were obtained by using the full spectrum. We specify the entropy per particle s = S/N, which we find to be approximately independent of coupling parameters in the experimental realizations. The results in the Hubbard regime are calculated using the DMRG ansatz^{44} based on the TeNPy library (v.0.3.0)^{45}. For all calculations, we conserved the total particle number and the total magnetization. For the phase diagram in Fig. 4d we used the iDMRG technique to obtain the ground state and the values of the string order parameters in the thermodynamic limit. For this, we evaluated the ground state for each parameter and used a maximal MPS bond dimension χ = 1,200. The bond dimension is increased in steps Δχ = 40 and the simulation stopped when the difference in the ground state energy per unit cell E(χ + Δχ) − E(χ) < 10^{−7}. This worked for most parameters except in the vicinity of two decoupled Hubbard chains and at small values of \(U/{t}_{\parallel }\). Nevertheless, in this regime we find that the energy per unit cell is converged up to E(1,200) − E(1,160) < 10^{−4}. For the experimentally accessible regime all calculations fulfil the former bound. To obtain the infinite length value of the string order parameters, we calculated it for different lengths \(d\in \)[200, 400, ..., 1600] to make sure that its final value is converged.
Data availability
The datasets generated and analysed during the current study are available from the corresponding author on reasonable request. Source data are provided with this paper.
Code availability
The code used for the analysis are available from the corresponding author on reasonable request.
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Acknowledgements
This work was supported by the Max Planck Society (MPG), the European Union (FETFlag 817482, PASQUANS), the Max Planck Harvard Research Center for Quantum Optics (MPHQ), the Cluster of Excellence ‘CUI: Advanced Imaging of Matter’ of the Deutsche Forschungsgemeinschaft (DFG) – EXC 2056 – project ID 390715994 and under Germany’s Excellence Strategy – EXC2111390814868. J.K. acknowledges funding from Hector Fellow Academy and T.C. from the Alexander v. Humboldt foundation. R.V. is supported by the Harvard Quantum Initiative Postdoctoral Fellowship in Science and Engineering and by the Simons Collaboration on UltraQuantum Matter (Simons Foundation, 651440, Ashvin Vishwanath). F.P. acknowledges the support of the European Research Council (ERC) under the European Union’s Horizon 2020 research and innovation program (grant agreement no. 771537). P.S. acknowledges support from the NSRF via the Program Management Unit for Human Resources & Institutional Development, Research and Innovation (grant number B05F640051).
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P.S., S.H. and D.B. planned the experiment and analysed the data. P.S., S.H., D.B., T.C. and P.B. collected the data. T.C. carried out the ED calculations. J.K. and G.S. contributed to the upgrade of the experimental apparatus. J.B., R.V. and F.P. performed the DMRG simulations and analytical calculations. T.A.H., C.G. and I.B. supervised the study. All authors contributed extensively to interpretation of the data and production of the manuscript.
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Extended data figures and tables
Extended Data Fig. 1 Density engineering.
a, Repulsive light shaped with a DMD splits the system into four independent ladders in the centre surrounded by a lowdensity bath. The density of the ladders is n = 0.992 with a standard deviation of 0.03. b, The occupation histograms show the normalized occurrence of total atom numbers in each ladder and the normalized occurrence in the surrounding bath for L = 7. Almost 25% of the ladder realizations have N = 2L.
Extended Data Fig. 2 Nearestneighbour spin correlations.
The nearestneighbour spin correlation C(1) for different \({J}_{\perp }/{J}_{\parallel }\) in the L = 5 system. The brown (purple) points refer to the correlations along the rung (leg). The shaded areas correspond to the correlations in the Heisenberg model with an entropy of S/N = (0.3−0.4) k_{B} per particle. Both theoretical and experimental values are obtained from the magnetization sector \({m}^{z}\) = 0.
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Sompet, P., Hirthe, S., Bourgund, D. et al. Realizing the symmetryprotected Haldane phase in Fermi–Hubbard ladders. Nature 606, 484–488 (2022). https://doi.org/10.1038/s4158602204688z
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DOI: https://doi.org/10.1038/s4158602204688z
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