Abstract
The Hubbard model is an essential tool for understanding manybody physics in condensed matter systems. Artificial lattices of dopants in silicon are a promising method for the analog quantum simulation of extended FermiHubbard Hamiltonians in the strong interaction regime. However, complex atombased device fabrication requirements have meant emulating a tunable twodimensional FermiHubbard Hamiltonian in silicon has not been achieved. Here, we fabricate 3 × 3 arrays of single/fewdopant quantum dots with finite disorder and demonstrate tuning of the electron ensemble using gates and probe the manybody states using quantum transport measurements. By controlling the lattice constants, we tune the hopping amplitude and longrange interactions and observe the finitesize analogue of a transition from metallic to Mott insulating behavior. We simulate thermally activated hopping and Hubbard band formation using increased temperatures. As atomically precise fabrication continues to improve, these results enable a new class of engineered artificial lattices to simulate interactive fermionic models.
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Introduction
Analog quantum simulators are designed quantum systems with a tunable Hamiltonian to emulate complex quantum systems intractable using classical computers due to the exponential growth of the Hilbert space with the system size^{1}. Simulating strongly interacting fermions on a lattice lies at the heart of understanding quantum manybody phenomena, such as highTc superconductivity^{2} and spin liquidity^{3} that emerge in solidstate systems at low temperatures and are not describable through meanfield or density functional theory.
Various experimental platforms that form artificial lattices have been explored for realizing FermiHubbard analog quantum simulators, including optical lattices^{4,5}, moiré superlattices^{6}, and semiconductor quantum dot systems^{7,8}. Quantum dots, often referred to as artificial atoms, can be arranged into artificial molecules and lattices with tunable hopping amplitude, interaction strength, and customdesigned point symmetry. For probing Fermionic manybody physics, the unique advantages of quantumdot systems relative to other platforms, such as cold atoms in optical lattices, include a readily achievable lowtemperaturelimit with respect to the hopping amplitude, easy access to transport measurements, and dynamic control of the chemical potential landscape and filling factors using gates^{9}. Amongst the various semiconductor quantum dot systems, lattices of dopantbased quantum dots have unique advantages in simulating strongly correlated Fermionic systems of real atomic lattice sites because the atomic nature of the quantum dots means they have naturally occurring ioncores, nuclear spins, hyperfine interactions, and inherently strong longrange interactions. Additionally, patterning the device geometry using the scanning tunneling microscope (STM)based hydrogen lithography technique^{10} adds the versatility of tailoring complex gate designs.
Effective control of tunable Hamiltonian parameters and precisionengineering of electron and spin correlations in a dopantbased artificial lattice relies on the controlled placement of dopant atoms in the host lattice with nearatomic precision. Although the AndersonMott transition has been previously demonstrated in fewatom systems using ionimplanted single dopant impurity atoms in silicon^{11}, this technique is limited by implantation aperture size and ion straggle. To our knowledge the best reported implant positioning accuracy is ~5–10 nm^{11}, which is incompatible with multiple site atomicprecision dopant arrays (In this study, we define “atomicscale precision” as achieving subnm positioning accuracy of the artificial lattice sites using single/fewdopant quantum dots). The STMbased hydrogen lithography technique, initially pioneered by Joseph Lyding’s group^{12}, was further developed and expanded by Michelle Simmons’ group at UNSW^{10,13} into a method that allows controlled placement of dopant atoms buried in an epitaxial Si environment, enabling a new suite of atomicscale precision devices. Several groups have further developed this atomicscale device fabrication technique introducing new applications, e.g., NIST (dopantbased analog quantum simulation)^{14,15}, Sandia (CMOS integration)^{16,17}, and UCL,^{18,19}. These advances have demonstrated success in atombyatom construction of single and fewdopant quantum dot devices in silicon and atomicscale control of tunneling in dopantbased devices, enabling highfidelity dopantbased multiqubits^{20} and simulation of the topological phases of the 1dimentional manybody Su–Schrieffer–Heeger (SSH) model^{21}. Here, we develop a path towards precise fabrication of dopantbased FermiHubbard simulators where the onsite electronelectron interactions can be controlled by engineering the number and configuration of dopant atoms at each site, while the sitetosite hopping amplitude and strength of longrange interactions can be controlled by altering the spatial separations between sites. Limited theoretical studies have been carried out to predict the manybody properties in dopantbased quantum dot arrays in silicon^{22,23,24}, such as strong correlation, excitation spectrum, and their robustness against disorder for analog quantum simulation. Transport through manybody states in dopantbased arrays has also been proposed as a sensitive probe to topological phase transitions^{24} and coherent manipulation of electronics states^{25}. Until now, however, atomicscale fabrication and quantum transport characterization of 2D artificial lattices of single/fewdopants have not been realized in the laboratory.
Here, we simulate an extended 2D FermiHubbard Hamiltonian using atomicscale fabricated 3 × 3 lattices of single/fewdopant quantum dot, although disorder present in the arrays leads to uncertainty in the underlying Hamiltonian. We define the electron ensemble in the array by tuning the chemical potential landscape using inplane gates and measure the lowtemperature quantum transport through the array to probe the charge addition spectrum, resonant tunneling, and the impact of inhomogeneity within the manybody system of the arrays. By increasing the average lattice constant from ~4.1 nm to ~10.7 nm, we tune the sitetosite hopping amplitude and longrange interactions within the simulated Hamiltonian and observe a transition from metallic behavior to the formation of Coulomb blockade. Finally, at elevated temperatures, we observe the formation of Hubbard bands in transport spectroscopy that can be attributed to additional hopping introduced by thermally activated occupation of manybody states. To augment the interpretation of our simulated extended Hubbard Hamiltonian, we numerically solve the Hubbard Hamiltonian using exact diagonalization and parameters estimated based on the experimental characterization of the array. Our results establish a new solidstate platform for the exploration of extended FermiHubbard models of strongly correlated 2D systems.
Results
3 × 3 arrays of fewdopant quantum dots and the extended FermiHubbard Hamiltonian
We fabricated^{14} a series of 3 × 3 square lattice arrays of fewdopant quantum dots that are weakly tunnel coupled to a source and drain and capacitively coupled to two inplane gates. Figure 1 shows STM images of the hydrogen lithography patterns from one of the arrays (average lattice constant \(a=10.7\pm 0.3{nm}\)) on a hydrogenterminated Si(100) 2 × 1 reconstruction surface, where the locations of the artificial lattice sites and the lattice constants can be determined by using the surface reconstruction dimer unit cells as an atomically precise ruler and counting the number of dimer rows (dimerrow pitch = 0.77 nm) between neighboring sites. Each lattice site is defined by using an STM tip to remove a small patch of (~10 to ~20) adjacent hydrogen atoms, allowing individual phosphorus atoms to incorporate only into the exposed surface Si lattice sites in a subsequent phosphorus dosing and incorporation process. We have recently demonstrated that a dangling bond patch of similar size typically forms a fewdopant cluster quantum dot incorporating 1 to 3 phosphorus atoms^{14}. At the same time, the STMpatterned inplane source/drain leads and two symmetric inplane gates are saturationdoped to a dopant density of ~\(2\times {10}^{14}/{{cm}}^{2}\) that corresponds to a bulk doping density of \(2\times {10}^{21}/{{cm}}^{3}\), approximately three orders of magnitude above the bulk metalinsulator transition, allowing quasimetallic conduction in all electrodes. We use a roomtemperature grown locking layer technique to suppress atomicscale movement of the precisely defined dopant atom positions^{26} before a subsequent lowtemperature (~250 °C) epitaxial Si overgrowth, that embeds the dopant atoms in a 3dimensional crystalline Si environment. Finally, ohmic contacts to the buried electrodes are formed using a low thermal budget silicide contacting technique^{27}.
The Hubbard model has long provided a theoretical playground for understanding different phases of matter, especially in the presence of strong electronic correlations. The 3 × 3 array’s physical attributes map to an extended FermiHubbard Hamiltonian, which, in its simplest form, includes one spinful orbital at each site. Experimentally, the absolute number of excess electrons at each single/fewdopant quantum dot may vary according to the inhomogeneity in the number of dopant atoms at each lattice site. For quantum transport through dopantbased quantum dots within the small bias ranges used in this study, charge fluctuations at quantum dots occur via quantum dot energy levels that are near or in between the Fermi levels in the source and drain leads. Therefore, we limit our analysis to onsite binding energy levels that are nearest to the Fermi level, i.e., charge number fluctuations of up to two electrons at each site which corresponds to the three charge states of a fewdopant quantum dot: the ionized state, the chargeneutral state, and the negatively charged state. The absolute number of excess electrons on a fewdopant quantum dot does not affect the underlying physical phenomena in this work.
The total Hamiltonian consists of the onsite energy terms in H_{μ}, the kinetic energy (hopping) terms in \({H}_{t},\) and the electronelectron interaction energy terms in \({H}_{U}\). Here, \({n}_{i\sigma }={c}_{i\sigma }^{{{\dagger}} }{c}_{i\sigma }\) is the number operator where \({c}_{i\sigma }^{{{\dagger}} }\) (\({c}_{i\sigma }\)) is the creation (annihilation) operator of a fermion with spin \(\sigma\) at lattice site \(i\). μ_{i} is the chemical potential at site \(i\) comprised of fixed contributions from local and longrange electronion core Coulomb interactions, i.e., \({E}_{{bi}}\) (electron binding energy at site \(i\)) and \({V}_{i,j}\) (Coulomb attraction between an electron at site \(i\) and an ioncore at site \(j\)); \({E}_{{bi}}\) is determined by the number of dopants and detailed dopantcluster configurations at each site. \({V}_{i,j}\) is determined by the separation between two lattice sites and can be approximated using a pointcharge approximation \({V}_{i,j}\approx \frac{{V}_{0}}{\left{R}_{i}{R}_{j}\right}\) where \({V}_{0}=\frac{{e}^{2}}{4\pi {\varepsilon }_{r}{\varepsilon }_{0}}\approx 123{meV}\cdot {nm}\) in silicon (see Supplementary Note 1). \({\mu }_{i}\) also includes a tunable contribution \({p}_{i}\) that is determined by the classical capacitance couplings of the device and the applied voltages on the gates and source/drain leads. \(t\) represents the hopping amplitude between nearestneighbor sites (\(H.c.\) indicates Hermitian conjugate). \({U}_{i}\) is the local electronelectron Coulomb repulsion at site \(i\), and \({U}_{{ij}}\) is the longrange Coulomb repulsion between electrons at sites \(i\) and \(j\). The filling factor (electron number) in the array is determined by the position of the array’s chemical potential with respect to the Fermi levels in the source/drain leads, which we set as \({E}_{F}=80{meV}\) below the Si conduction band edge^{10}. We ignore Coulomb exchange and higherorder hopping terms in our numerical model^{28}.
The hopping \(t\), interactions \({U}_{i}\), \({U}_{i,j}\), chemical potential terms μ_{i}, which determines the electron distribution and doping level of the array, and the temperature \(T\) constitute the FermiHubbard model’s parameter space, covering a large variety of correlated electron phenomena, some not yet fully understood, and others that continue to be discovered. Physical control of the Hubbard model parameters in our 2D arrays is achieved by varying device fabrication and measurement conditions. The number of electrons in the array (\(\mathop{\sum}\limits_{i}{n}_{i}\)), can be altered by applying a common voltage on both inplane gates to shift the array’s chemical potential landscape (albeit nonuniformly due to screening from source and drain leads) with respect to the Fermi level. A voltage difference between the two gates will introduce a potential gradient within the 2D array along the gategate direction. In contrast to gatedefined quantum dot systems, insitu gate tuning of tunnel coupling within a single device is less efficient in donordefined quantum dot systems. Effects of tuning the hopping amplitude \(t\) and longrange interaction terms (\({U}_{i,j}\)) can be achieved, however, by fabricating a series of dopantbased lattices with different lattice constants that are determined at the fabrication stage. The local interaction term \({U}_{i}\) reflects the physical size of each quantum dot. For the arrays in this study, we design the inplane gates and average lithographic dot size to be the same for all arrays and only alter the lattice constant from \({a}_{1}=4.1\pm 0.3{nm}\) in the first array to \({a}_{2}=6.6\pm 0.3{nm}\) in the second array and to \({a}_{3}=10.7\pm 0.3{nm}\) in the third array. Based on previous theoretical studies^{28,29}, these lattice constants correspond to hopping amplitudes in the ~8 \({meV}\) to the hundreds of \(\mu {eV}\) range and longrange interactions in the range of ~20 meV to few\({meV}\) (See Supplementary Note 1). We estimate the number of dopant atoms in the quantum dots to be within the range of \(2\pm 1\) dopants by characterizing the binding energies of fewdopant quantum dots with a similar lithographic patch (See Supplementary Notes 2 and 3). A quantum dot size of 2 ± 1 dopants corresponds to a local electronelectron interaction energy \({U}_{i}\) of ~45 meV (for the relevant chargeneutral to negatively charged state transitions). These energy scales position these Hubbard arrays in the strongly interacting regime with nonnegligible longrange interactions, even beyond nearestneighbor sites^{30}. Our best estimates point to an average ratio \({U}_{i}/t\) that varies roughly from 6 to 90 from the first to the third array. The ratio of the nearestneighbor Coulomb repulsion and the hopping amplitude also varies by an order of magnitude roughly from 2 to 20. So, while we expect Coulomb interactions to dominate in the third array, their strengths are comparable to, or less than, the noninteracting bandwidth (see SM, Fig. S4) in the metallic array, and therefore, we expect the significant tunneling/delocalization of electrons to change the character of the latter system. Disorder also plays a significant role in all three arrays. Apart from introducing sitetosite variations in interactions and tunneling amplitudes, the ratio of its strength (defined as the difference in binding energies for 1 and 3 dopants per site) and the hopping amplitudes varies roughly from 4 to 70 from the first to the third array.
Atomicscale defects have been a critical challenge for solidstate implementations of quantum devices that rely on atomic precision fabrication processes such as those used here. The primary source of defect/disorder is the sitebysite variation in onsite energies \({E}_{{bi}}\) and \({U}_{i}\), which effectively introduces inhomogeneity in the chemical potential landscape, interactions, and hopping amplitudes. Due to the stochastic nature of the phosphorus dosing and incorporation process^{31}, deterministic control of the exact number of dopant atoms and their specific cluster configuration within a lithographic patch remains an unsolved challenge in the community^{16}. While the precise atomic configurations in an array may, in principle, be obtained by parametrically fitting the measurement results with numerical simulations, a detailed disorder configuration investigation is extremely computationally expensive and beyond the scope of this study. Instead, we account for disorder by estimating the number of dopant atomspersite, based on STMlithography patterns and dopant incorporation conditions^{32}, for use as input to the numerical simulations. We do not distinguish the effects from different sources of disorder because variation in dopant number per site and their clustering configuration have similar effects on sitebysite energy variations. While we have not pursued the exact match between the numerically simulated and experimentally simulated results in this study, the detailed atomic configuration at each array site does not alter the qualitative understanding of the array system (see Supplementary Note 6), and the quantitative differences between theory and experiment are a measure of the accuracy of the disorder estimates. Supplementary Table 3 in the Supplementary Notes lists the range of binding energy \({E}_{b}\) and the onsite addition energy \({U}_{i}\) for 1 P, 2 P, and 3 P dopant clusters^{29}. Variations in the number of dopant atoms and detailed dopant cluster configurations at each lattice site introduce site disorder, which in larger arrays lends itself to the study of Anderson/Mott localization, and especially its existence in the presence of strong interactions^{27}.
In the following sections, we demonstrate the tunability of dopantbased 2D arrays by first showing tuning of the hopping amplitudes and longrange interactions through transport measurements of arrays fabricated with different lattice constants. We then demonstrate the ability to define the electron ensemble (the number of electrons in the manybody system of the array) and characterize the charge addition boundaries and resonant tunneling within the 2D array using inplane gates. Finally, we measure the transport spectrum at elevated temperatures and reveal thermally activated Hubbard bands within the array.
Tuning the hopping amplitude and longrange interactions
It is well known^{28} that the nearest neighbor hopping \(t\) is exponentially dependent on the latticeconstant \(a\) while the longrange ee Coulomb interactions are inversely proportional to \(a\). Figure 2a, d, g shows STM images after patterning from three arrays, where we intentionally increase the lattice constant \({a}_{1}=4.1\pm 0.3{nm}\) in the first array, to \({a}_{2}=6.6\pm 0.3{nm}\) in the second array, and to \({a}_{3}=10.7\pm 0.3{nm}\) in the third array. If we interpret each artificial lattice site as an ‘impurity’ atom in silicon, the lattice constant in the first, second, and the third arrays correspond to a bulk doping density of \(\sim 1.5\times {10}^{19}/{{cm}}^{3}\), \(\sim 3.5\times {10}^{18}/{{cm}}^{3}\), and \(\sim 8\times {10}^{17}/{{cm}}^{3}\), spanning from above to below the critical density of a metalinsulator transition in phosphorusdoped bulk silicon. As will be described in the following, transport measurement and charge stability analysis confirm a transition from delocalized electrons displaying metallic behavior in the first array to localized electrons in the third array.
Figure 2b, e, h are the measured charge stability diagrams from the three arrays measured at T ~ 10 mK base temperature (electron temperature ~300 mK). In the first array, charge addition boundaries run more or less parallel throughout the entire gategate plane, resembling the charge stability diagram of a single metallic island in a singleelectron transistor (SET). Comparing the charge stability diagrams of the first array with the second and third arrays, a key observation is that the charge stability boundaries in the third array are dominated by straight line segments of distinct negative slopes with bias triangles at avoided crossings; whereas on the gategate plane in the second array, the charge addition boundaries appear crossed only at negative gate voltages and evolve into smooth curves at positive gate voltages. This is further substantiated by comparing the differential conductance bias spectra of the first and second arrays (Fig. 2c, f) measured along the dashed lines in Fig. 2b, e, respectively. Here, the Coulomb diamonds in the first array (Fig. 2c) are dominated by welldefined diamond shapes that are closing at small biases and of similar diamond heights (charging energy), resembling the conventional Coulomb blockade behavior of a metallic island SET. However, despite the dominant metallic behavior in the gategate map, we do observe small discontinuities in the transport lines in the higher voltage range, likely due to nearby charge instabilities.
In contrast, the Coulomb diamonds of the second array (Fig. 2f) are irregularly shaped, and, despite the expected presence of disorder within the arrays, the bias spectrum of the second array features two groups of small Coulomb diamonds (highlighted by magenta color bars), corresponding to the upper and lower Hubbard bands, separated by a set of large, irregularly shaped Coulomb diamonds, corresponding to the finite analog of a Mott gap. The separation of the two Hubbard bands is characterized by the heights (addition energy) of the Coulomb diamonds at the Mott gap, ~50 meV, which is in reasonable agreement with the addition energy of individual quantum dots of few (1 to 3) dopant atoms^{29}. (See Supplementary Note 3) Additionally, in Fig. 2c, f, we observe a qualitative difference in the bias conductance spectrum between the two arrays at base temperature. In the second array (Fig. 2f), conductance is visible as lines of increased differential conductance running parallel to the edges of Coulomb diamonds, indicating a discrete eigenenergy spectrum within the array; such a discrete conductance spectrum is not visible in the first array (Fig. 2c), indicating a quasicontinuous (metallic) density of state distribution in the first array, i.e., the exited addition energy state separation falls below \({k}_{B}T\) so the addition energy levels can no longer be individually distinguished at the operating temperature. (See Supplementary Note 4 for the impact of hopping, longrange interactions, and disorder on simulated addition energy spectrum and charge stability diagrams and Supplementary Note 5 for the effects from decreasing the lattice constants within the simulated array)^{33}.
Gatetuning the electron ensemble and charge distributions
In this section, we use results from the third array, which is near the atomiclimit (weak tunnel coupling limit, \(\frac{U}{t}\gg 1\)) to illustrate tuning of the charge distributions within an array using inplane gates. In Array #3, since the source/drain leads are weakly tunnel coupled to the array and there is relatively large local and longrange electronelectron interactions, the sourcedrain conductance through the array is in the Coulomb blockade transport regime^{34,35}, as evidenced by the zeroconductance regions at finite bias in the measured conductance maps. The Coulomb blockade in Array #3 can be lifted by applying plunger gate voltages that overcome the interactioninduced blockade barrier and align an addition energy level of the array with the Fermi level in the source and drain leads. At these gate conditions the electron number in the array can fluctuate by one (finite compressibility), allowing sourcedrain conductance through the array.
Figure 3a, b plot numerically simulated charge stability diagrams of the ground states and conductance map over the gategate space for Array #3. Within each color domain in the simulated charge stability diagram, the total number of excess charges in the array (\(N\)) is constant, corresponding to the Coulomb blockaded regions in transport. Sweeping the commonvoltage applied to both gates along the 45degree diagonal direction on the charge stability diagram controls the total number of electrons, therefore, the filling factor of the model. Sweeping the differential voltage between the two inplane gates (along the 135degree diagonal direction in Fig. 3a) effectively tilts the chemical potential landscape within the array, altering the charge distribution of the ground state without considerably affecting the filling factor within the array. Due to variations in the input binding energies (see Supplementary Note 2), the largest Coulomb blockaded region in the theory diagram belongs to N = 8, corresponding to ~11% hole doping, as opposed to half filling (N = 9), which is expected for the uniform FermiHubbard model.
The charge stability domain boundaries correspond to conductance lines in a transport map. Comparing the simulated conductance map with the charge stability diagram, not all charge addition boundaries are equally visible in conductance. This is also evident in the nonclosing diamonds that are circled in Fig. 2i^{36}. Both the measured and simulated conductance maps agree on a relatively large conductance along a particular line of negative slope in the positive gate voltages region of the diagram (circled in Figs. 2h and 3b). Visibility of the charge addition boundaries in measured conductance maps can be enhanced by increasing the bias window (see Supplementary Note 8) to allow transport through excited states or other inelastic/incoherent transport processes in the array mediated by, for example, electronphonon interactions. We attribute the nonsymmetric shape of the charge stability diagrams with respect to two gate voltages, or the fact that Coulomb diamonds do not always close (do not have conductance) at a small bias along the diagonal gategate direction, to the following: First, we expect the dominant effect to be the nonuniformity in binding energies due to variations in the number of dopants per site, which not only makes the system’s energy landscape nonuniform for the electrons but also affects the transport through the array. Additional effects include small asymmetries in the way each gate affects individual site potentials and atomicscale variation in the hopping amplitudes.
Effects of reduced lattice constants and increased hopping amplitudes for the first and the second arrays, where t = 8 meV and t = 2 meV, respectively, can be seen in simulated charge stability and conductance diagrams (see Supplementary Section S8 and Fig. S9). To illustrate the delocalization of manybody states in the second array, Fig. S9c plots ground state charge and charge addition distributions at select positions near an avoided crossing. Figure 3c shows ground state charge and charge addition distributions at selected positions near an avoided crossing at \(N=8\) in the simulated charge stability diagram in Fig. 3a. The simulations confirm that electron additions take place mostly locally at individual sites on this array. Comparisons to similar plots for the second array (see Supplementary Note 5) reveal the effect of reduced lattice constants and increased hopping amplitudes (t = 2 meV) in delocalizing the added electrons.
With our symmetric, inplane twogate design, the ratio of the upper gate’s (Gate1) and lower gate’s (Gate2) capacitive coupling to electrons at a quantum dot is determined by the ratio of the quantum dot’s distances to the upper gate and to the lower gate. Charge additions to quantum dots in the upper, middle, and lower rows for Array #3 correspond to three distinct lever arm ratios, which manifest in the charge stability diagram as three different slopes in charge addition boundaries. For example, in Fig. 3c, when gating the array from positions 1 to 2 and from 3 to 4, we cross charge addition boundaries of the same slope, corresponding to the addition of an electron onto sites in the lower row. Similarly, when gating the array from positions 1 to 3 and from 2 to 4, the charge addition boundary slope corresponds to adding an electron onto sites in the upper row. Charge addition to the middle row can be found elsewhere in Fig. 3a when crossing addition boundaries with an intermediate magnitude in slope.
We carry out a quantitative analysis of slopes by running edge detection algorithms on the measured charge stability maps of the three arrays (See Fig. 4a–c) and plotting histograms of the detected edge length along their negative slopes \(\frac{{\triangle V}_{G2}}{{\triangle V}_{G1}}\). As can be seen in Fig. 4d, e, f, the histogram of the first array (Fig. 4d), shows a single peak in the distribution, mimicking the behavior of conductance through a metallic SET island. In the second array (Fig. 4e), in contrast, the single peak in the distribution evolves into a broad distribution, exhibiting conductance that occurs via single addition electrons that are more often shared simultaneously by sites across different rows. The third array shows three narrow distinct peaks, corresponding to charge addition sites in the upper, middle, and lower rows, with little mixing in states between different rows for conductance via single electrons jumping on and off the array. This change in slope distributions is characteristic of the transition from metallic to the collective Coulomb blockade regime upon decreased hopping amplitude relative to the interaction strength^{34}. Such a transition from metallic behavior to the collective Coulomb blockade regime, where the energy gap inhibits adding an electron to an already strongly correlated state, is also observed in Fig. 2c, f as the manybody energy spectrum transitions from quasicontinuous to discrete distributions from the first to the second array. The extended skewing in slope distribution towards more negative slopes (\(\triangle {V}_{G2}/\triangle {V}_{G1}\)<−3) is likely the result of imperfect symmetry in device geometry and gate nonlinearity. Similar results are found in Fig. 4g, h, i plotting the distributions of the lever arm ratios of charge additions, which are calculated from the simulated charge stability diagrams using the identical bias windows as in the experimental plots to include charge addition configurations of exited manybody states. (see Supplementary Note 1 for detailed capacitance and lever arm calculation methods). We note that the agreement in the shapes of the slope distributions remains qualitative due to unknown disorder and real measurement conditions that are not fully accounted for in theory.
Resonant tunneling through manybody states
Additionally, we find small resonant tunneling lines with positive slopes in the conductance map from Array #3 (see Fig. 5a), which are attributed to hybridization between manybody states with the same filling factor but different charge distributions on the array. Because of the “multiply connected” topology^{37} in the 2D array, the resonance conditions within the array enhance conductance through the array by opening additional conduction paths. Figure 5b shows a histogram of the positive slopes of conductance resonance lines extracted from the charge stability diagrams of Array #3 over a bias range from +10 mV to −10 mV. An example of the bias evolution of a positively sloped resonance conductance region is shown in Fig. 5c. We attribute the difference in bias triangles at positive and negative biases to disorder in the array. From the histogram in Fig. 5b, we observe positive slope distributions centering at three distinct values representing resonant tunneling between manybody states that are primarily localized at the first and the second rows, at the first and the third rows, and at the second and the third rows of the quantum dot array, respectively. Examples are ground states at positions 2 and 3 in Fig. 3c with an avoided crossing of charge addition lines between them. The slopes of the two crossing charge addition lines (dashed lines in Fig. 5d) correspond to charge addition to sites in the upper row and to sites in the lower row, respectively. Approximating the two ground states as classical states of a doubledot system in which the electron is localized at either the left or the right dot, we can find an estimate for the effective tunneling between the states, t′, by fitting the conductance peak at the avoided crossing to the following Lorentzian form^{38,39}
where \({I}_{{ds}}\) is the drainsource current; \(t^{\prime}\) is the tunnel coupling rate (\(1{GHz}\,\approx \,4.14\mu {eV}\)) between two charge occupations of discrete energy levels. \({\varGamma }_{d}\) and \({\varGamma }_{s}\) are tunneling rates to or from the drain and source leads, which are assumed to be approximately equal in this study; \(\left(\varepsilon {\varepsilon }_{0}\right)\) is the detuning energy between the two discrete energy levels (see black dashed arrow in Fig. 5d, based on \({V}_{G1}\) and \({V}_{G2}\) using lever arms, see Supplementary Note 1); and \(h\) and \(e\) are Plank’s constant and the charge of a single electron. The amplitude of the current peak is primarily determined by \({\varGamma }_{d}\) and \({\varGamma }_{s}\) ; and the width of the peak depends mostly on \(t^{\prime}\). We treat \(t^{\prime}\), \({\varGamma }_{d}\), and \({\varGamma }_{s}\) as fitting parameters.
As shown in Fig. 5d, e, we perform the fit using the measured conductance near a crossing along the black dashed line in Fig. 5d. We find \(t^{\prime}\,\approx \,4.4\pm 0.3{GHz}\left(14.1\pm 1.7\mu {eV}\right)\) and \({\varGamma }_{d}\,\approx \,{\varGamma }_{s}\,\approx \,0.6\pm 0.1{GHz}\). The insensitivity of the resonant current peak height and peak shape to the applied bias confirms that the observed peaks are dominated by resonant transport through two discrete energy levels. Away from zerodetuning, the two discrete energy levels are isolated, while at zerodetuning, the two discrete energy levels maximally hybridize. For a doubledot system, the separation at an avoided crossing is determined by \(\left({U}_{m}+2t^{\prime}\right)\), where \({U}_{m}\) is the longrange ee repulsion between the two sites. We computationally verify that this relationship also applies to our 3 × 3 array systems where two localized charge distribution configurations that are on resonance at an avoided crossing can be mapped to the two sites of a doubledot system. (See Supplementary Note 8) Here each t’ should be interpreted as an offdiagonal matrix element for an effective twolevel Hamiltonian near the avoided crossing and distinguished from the nearest neighbor hopping amplitude t in the FermiHubbard Hamiltonian. However, we find that the magnitude of \({t}^{{\prime} }\) from the above fit is consistent with a simulated tunnel coupling^{40} that is between nonneighboring array sites in the upper and lower rows.
Thermally activated transport within an array
Transport through the second array is of special interest because its lattice site density corresponds to the critical doping density in silicon between tunneling and band transport regimes, where transitions of the electron system from a frozen Wignerlike phase to a Fermi glass has been previously observed by increasing the temperature in a fewdopant silicon transistor^{41}. Here, we explore the thermal activation in the second array by monitoring the temperature evolution of transport properties through the array. The experimental temperature range investigated is limited by the temperature (~16 K) at which leakage current through the Si substrate becomes comparable to conductance through the array. At elevated temperatures, the charge occupation is thermally broadened not only in the Fermi distributions within the source/drain leads, but also in the occupation of the manybody states within the array. In addition to reducing the opacity at the source and drain tunnel junctions, both thermal effects increase the ability for addition electrons to access manybody excited states, and therefore, open new tunneling paths within the array that were previously not accessible at base temperature.
Figure 6a plots the Coulomb oscillation in sourcedrain conductance at T = 10 mK and 12 K taken along the dashed line cut in Fig. 2e using a small bias voltage of \(3{mV}\). We observe the removal of the Coulomb blockade within each of the Hubbard bands at higher temperatures, where charge stability domains characterizing fixed numbers of electrons in the array become illdefined except for at the large finite analog Mott gap. We also observe a similar transition in transport in the simulated Coulomb oscillations at elevated temperatures (Fig. 6b), which can be explained by a transition from the Coulomb blockade regime to the collective Coulomb blockade regime (two Hubbard bands)^{34}. In this system, thermal broadening of Coulomb oscillation peaks at T = 12 K becomes large enough to eliminate smaller Coulomb blockade gaps within the upper and lower Hubbard bands but not the large gap in between. Such thermally activated Hubbard band formation behavior within a 2D array is a key distinction from the thermal broadening in a zerodimensional system where Hubbard bands do not exist.”
Figure 6c shows the temperature evolution in conductance at an avoided crossing region of the second array (circled in Fig. 2e) at T = 10 mK, 4 K, 8 K, and 12 K. Conductance at the two triplepoints at the avoided crossing stretches from welldefined bias triangles to conductance lines running in parallel, in analogy to the development of crescents at the triplepoints of a doubledot system when increasing the tunnelcoupling between the dots^{42}. Within the 2D array, this signature can be attributed to the additional hopping channels that become accessible at elevated temperatures (thermally activated hopping), that not only enhances conductance through the array, but also causes manybody states to become more delocalized within the array.
Discussion
We have simulated an extended FermiHubbard Hamiltonian of a 3 × 3 2D lattice using a series of STMfabricated 3 × 3 arrays of fewdopant quantum dots and probed the manybody properties within the arrays using quantum transport measurements. By introducing two inplane gates on either side of the array, we demonstrate gate tuning of the electron ensemble within the array as well as the flexibility of tilting the chemical potential landscape for characterizing the charge distributions of manybody ground states and resonant properties between them. By varying the lattice constants of a series of array devices at the fabrication stage, we demonstrate effective tuning of the hopping amplitude and longrange interactions within the simulated Hamiltonian. Through comparisons with theory, we can identify a finitesize analogue of the transition from metallic to Mott insulating behavior in charge distributions within the array when the lattice constant is increased from \(4.1\pm 0.3{nm}\) to \(10.7\pm 0.3{nm}\). By increasing the temperature near a critical lattice site density, we observe the formation of thermally activated Hubbard bands and enhanced electron transport through the array via thermally activated occupation.
In this first experimental realization of an extended FermiHubbard Hamiltonian using atomically patterned 2D arrays of dopants, the quantum simulation accuracy is limited by uncertainties in the exact number of dopant atoms and their clustering configurations within each array site, as well as atomicscale variations in nearestneighbor hopping amplitudes. Continuing efforts are underway within our group and in the research community to seek long term solutions in improving the atomic perfection in dopantbased quantum devices. In a separate study^{43}, we recently demonstrated fabrication of a sample having 6 sites and 12 sites with precisely one single dopant per site by combining lowtemperature feedbackcontrolled hydrogen lithography^{44} with a new method called feedbackcontrolled manipulation. The precise number of dopant atoms in a fewdopant quantum dot and its clustering configuration can be measured using STM spectroscopy after dopant incorporation and before Si epitaxial overgrowth. When combined with abinitio calculations of the electronic structure of STMmeasured dopantcluster configurations as inputs, we expect significant improvement in the agreement between the experiment and theory in the future. In addition, development of more complex virtual gate designs^{7} is underway allowing control of the chemical potential at each array site and fine tuning of local tunnel coupling, improving the accuracy of dopantbased analog quantum simulators in the 2 by n configuration.
These experiments confirm the viability of simulating a FermiHubbard model using dopantbased 2D arrays that account for limited nonuniformity and pave the way towards more complex and accurate analog quantum simulations of extended FermiHubbard Hamiltonians using dopant atoms. Tunability within the 2D array demonstrates the expected control of Coulomb interactions, hopping, filling factors, and temperature; this provides a pathway to explore previously inaccessible regions in the multidimensional condensed matter phase space. Extending this work to larger dopantbased arrays should allow the study of manybody localization and competing magnetic and charge order on 2D lattices, including frustrated geometries. In contrast to other quantum simulation platforms of FermiHubbard Hamiltonians such as optical lattices, dopantbased artificial lattices are unique in reaching low effective temperatures and easy access to quantum transport. Additionally, relative to gatedefined quantum dot arrays, a dopantbased system has the unique property of having the naturally occurring nucleus at each dopantbased lattice site, allowing simulations that include nuclear spins and hyperfine interactions inherent in realworld condensed matter physics.
With continued advances to reduce disorder and improve atomicscale precision in singledopant placement, this method can be generalized to larger lattice arrays and other types of dopant species, such as boron^{45} and arsenic^{19}, potentially embedded in alternative substrate lattice environments, such as a Ge substrate^{46}. The results demonstrated in this study serve as a launching point for a new class of engineered artificial lattice systems to explore strongly correlated manybody systems in the solid state and explore less wellunderstood phenomena such as the pseudogap phase, strange metals, topological phases, and superconductivity in the FermiHubbard model and magnetic ordering and frustration in spin systems. We point out that the problem of solving the Hubbard Hamiltonian for generic parameters quickly becomes intractable with stateoftheart classical computers (even for ground state properties) for systems larger than 5 × 5 lattice sites. Even the 3 × 3 arrays in this study are on the cusp of what can be numerically done today exactly for finitetemperature transport properties.
Methods
Device fabrication and measurements
The 3 × 3 fewdopant quantum dot arrays are fabricated in an ultrahigh vacuum (UHV) environment with a base pressure below \(4\times {10}^{9}\) Pa (\(3\times {10}^{11}\) Torr) using an STM tip to create atomicscale lithographic patterns of the device by removing individual hydrogen atoms on an hydrogenterminated, Si(100) 2 × 1 reconstructed surface^{44}. Detailed sample preparation, UHV sample cleaning, hydrogenresist formation, and STM tip fabrication and cleaning procedures have been published elsewhere^{47,48}. The substrate is then saturationdosed with PH_{3} gas at room temperature, where PH_{3} molecules selectively absorb only onto the lithographic regions where chemically reactive Si dangling bonds are exposed. A subsequent rapid thermal anneal process at 350 °C for 1 min incorporates the absorbed phosphorus atoms into the Si surface lattice sites while preserving the hydrogen resist to confine dopants within the patterned regions. We embed the incorporated dopant atoms in a single crystalline silicon environment using lowtemperature Si epitaxial overgrowth with an optimized locking layer to suppress dopant movement at the atomic scale during epitaxial overgrowth^{26,48}. Finally, ohmic contacts to the buried device are formed using a low thermal budget contacting technique^{27}. The number of incorporated P atoms in each quantum dot can be estimated by analyzing the lithographic sites that are available for P incorporation^{31}. See Supplementary Note 2 for analyzing the best estimates of dopant numbers at each lattice site. Using the Si(100) 2 × 1 surface reconstruction lattice as an atomically precise ruler, the lattice constant of the fabricated arrays (centertocenter distance between adjacent quantum dots) can be precisely characterized. The STMpatterned inplane source/drain leads and gates are saturationdoped^{49} with a dopant density of \(\sim \!2\times {10}^{14}/{{cm}}^{2}\), corresponding to a 3D doping density of \(\sim \!2\times {10}^{21}/{{cm}}^{3}\) that is approximately three orders of magnitude above the metalinsulator transition, allowing quasimetallic conduction in all electrodes. We carry out directcurrent (DC) transport measurement of the device inside a dilution refrigerator at a base temperature of ~10 mK.
Numerical simulation
We have constructed a timeindependent extended Hubbard model to simulate the 3 × 3 quantum dot array system following the analysis first introduced by ref. 50 for semiconductor quantum dots and more recently by Le and coworkers^{28} for dopant arrays in silicon. In our extended Hubbard model (see Eq. 1), we include one spinful orbital per dot near the Fermi level, which means a maximum of 18 valence electrons can be added onto the array. We use a finite element software package, FasterCAP^{51}, for capacitance matrix calculations based on STM imaging of device geometry patterned using STMlithography, where a lateral seam of 2.5 nm and a vertical thickness of 2 nm have been included to account for the Bohrradiuslike electron density extension beyond the lithographic patterns, previously shown to be necessary to reproduce the experimental capacitance values in STMpatterned Si:P devices^{15,52}. We formulate the electrostatic potential landscape as well as onsite and longrange electronelectron Coulomb interactions using the calculated capacitance matrix (See Supplementary Note 1). We estimate the number of dopant atoms in the quantum dots by characterizing the binding energies of similar fewdopant quantum dots with a similar lithographic patch (See Supplementary Note 3). We approximate the screened onsite electronion Coulomb interaction in a quantum dot using the chargeneutral binding energy of the estimated dopant cluster^{29}. We adopt a Fermi level of −80 meV with respect to Si conduction band edge in the saturation doped, inplane source/drain, and gate electrodes^{53,54}. We approximate the longrange electronion Coulomb attractions using a point charge approximation^{28}. More details about model parameter extractions are provided in the Supplementary Notes. Exchange interactions and longerrange hopping terms are not implemented in our model. We assume the lead’s tunnel coupling to the dot array is sufficiently weak so that the dot array can be treated as an isolated system when solving for the eigenstates.
At each gategate point, we use exact diagonalization to solve for the ground and excited manybody eigenstates of the Hamiltonian, whose matrix is represented in the Fock basis, for particle numbers ranging from 0 to 18 in the 3 × 3 array system. The charge stability diagram simulations utilize the classical capacitance coupling and bestestimated numbers of dopants per site based on the STMpatterned device geometries (See Supplementary Note 2). The hopping amplitudes are set as t = 8 meV, t = 2 meV, and t = 0.5 meV for simulating the first, second, and third arrays, respectively, based on the designed lattice constants in the arrays^{40}. Charge stability diagrams and ground state charge distributions are obtained using the Lanczos algorithm after further blockdiagonalizing the Hamiltonian matrices based on the total spin in the z direction. The numerically simulated conductance results in the main text are linear response conductance through the array at finite temperatures and zero bias that is calculated following the formalism of ref. 28, which uses Fermi’s golden rule for the tunneling rate to/from the leads and a generalization of the transport equations developed by Beenakker^{55} for a single quantum dot. For that, we implement full reorthogonalization of vectors in the Krylov space in our Lanczos routine after each iteration and seek convergence for 25 to ~40 lowlying states if the size of the Hilbert space for a particular spin and particle number sector is greater than 2000. For that, except in Fig. 6b where we have used full diagonalization to obtain exact results, this approximation generally underestimates the conductance values, especially at elevated temperatures (T > 1 K), but does not alter its qualitative behavior. For smaller sizes, we perform full diagonalization using Intel’s math kernel library. The experimental conductance gategate maps were measured at finite bias voltage of the order of few mV and the extent to which excited manybody states participate in transport is largely determined by the bias window, rather than thermal broadening. The calculated conductance does not take into account the effect of the environment, including the leads and the bias window, or other processes, such as phonons, which can further broaden the measured conductance peaks even at very low temperatures. Therefore, we approximate these effects and the participation of excited states in transport by setting a 1 meV thermal broadening in the numerical calculations of conductance gategate maps of Fig. 3b in the main text and Supplementary Fig. 5b in the Supplementary Notes.
Adopting a similar notation as ref. 28, we write the conductance as
Where \({g}_{T}={e}^{2}\Gamma /\left({{\hslash }} {kT}\right)\),
and
where the last sum runs over sites on the left edge of the array, closest to the source lead. \({M}_{\alpha,\beta }^{\left(R\right)n,m}\) has a similar formula with a sum that runs over sites on the right edge of the array. \({E}_{\alpha }^{n,m}\) is the \({\alpha }^{{th}}\)eigenenergy of the manybody wavefunction, \({\Psi }_{\alpha }^{n,m}\) with \(n\) particles and a total spin of \(m\). \(k\) is Boltzmann’s constant\(,\) and \({f}_{{FD}}\) is the Fermi–Dirac distribution function. \(\Gamma\) is the hopping amplitude for electrons to/from the leads. In our plots, we show results for the dimensionless quantity \(G/{g}_{T}\). We use spin inversion symmetry to simplify our calculations by considering the transport of only spinup electrons through the array and inserting an overall factor of two in \(G\) (Certain commercial equipment, instruments, or materials are identified in this paper to foster understanding. Such identification does not imply recommendation or endorsement by the national institute of standards and technology, nor does it imply that the materials or equipment identified are necessarily the best available for the purpose.).
Data availability
All relevant data are available upon request from the authors.
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Acknowledgements
We acknowledge helpful discussions with John Randall and James Owen as well as Emily Townsend, Neil Zimmerman, Alexey Gorshkov, and Maicol A. Ochoa. This research was funded in part by the Department of Energy Advanced Manufacturing Office Award Number DEEE0008311 and by a National Institute of Standards and Technology (NIST) Innovations in Measurement Science award, “AtomBased Devices: Single Atom Transistors to Solid State Quantum Computing.” E.K. was supported by the National Science Foundation under Grant No. DMR 1918572. This work was performed in part at the Center for Nanoscale Science and Technology NanoFab at NIST.
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X.W. performed the measurements, X.W. and R.S. analyzed the data, X.W., J.W., and P.N. fabricated the devices. X.W. and E.K. formulated the theoretical model and carried out the numerical simulations. X.W., R.S., E.K., F.F., and G.B. contributed to the theoretical interpretation of the data. X.W., A.R., R.K., F.F., and R.S. contributed to electrical measurements. X.W., E.K., and R.S. wrote the manuscript with comments from all coauthors. R.S. conceived and supervised the project.
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Wang, X., Khatami, E., Fei, F. et al. Experimental realization of an extended FermiHubbard model using a 2D lattice of dopantbased quantum dots. Nat Commun 13, 6824 (2022). https://doi.org/10.1038/s4146702234220w
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DOI: https://doi.org/10.1038/s4146702234220w
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