Abstract
In a thin Weyl semimetal, a thickness dependent Weylorbit quantum oscillation was proposed to exist, originating from a nonlocal cyclotron orbit via electron tunnelings between top and bottom Fermiarc surface states. Here, magnetotransport measurements were carried out on untwinned Weyl metal SrRuO_{3} thin films. In particular, quantum oscillations with a frequency F_{s1} ≈ 30 T were identified, corresponding to a small Fermi pocket with a light effective mass. Its oscillation amplitude appears to be at maximum for thicknesses in a range of 10 to 20 nm, and the phase of oscillation exhibits a systematic change with film thickness. The constructed Landau fan diagram shows an unusual concave downward curvature in the 1/μ_{0}H_{n}n curve, where n is the Landau level index. From thickness and fieldorientation dependence, the F_{s1} oscillation is attributed to be of surface origin. Those findings can be understood within the framework of the Weylorbit quantum oscillation effect with nonadiabatic corrections.
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Introduction
The phenomena of quantum oscillations with magnetic field in matters were first discovered in bismuth more than 90 years ago^{1}, and it was soon realized that such phenomena can be used as a measure for the Fermi surface in momentum space^{1,2}. Quantitatively, quantum oscillatory phenomena are described by the general expression of Δρ ∝ Dcos[\(2\pi (\frac{F}{{\mu }_{0}H}+\delta )\)], where D is the damping factor, F is the oscillation frequency, and δ is the phase shift of the oscillation associated with individual extremal areas of the Fermi surface in momentum space. In theory, the absolute value of δ is a consequence of the underlying band structure and dimensionality of the material under study in which the Berry phase is derived as 0 or π. From a modern point of view, the periodic oscillation of either ρ or magnetization with respect to the inverse of a magnetic field is a quantum effect deriving from the BohrSommerfeld quantization rule on electrons’ cyclotron motions, where the requirement of the singlevalued wave function gives rise to discrete energies for electrons^{3}. The oscillation frequency (F) can be linked directly to the extremal crosssectional area (A_{e}) of the Fermi surface by Onsager’s relation of F = ℏA_{e}/2πe. In addition, the amplitude and phase of quantum oscillations encode the information about the electronic band topology at the Fermi surface^{2} that has been a major subject recently in the field of topological materials^{4,5}.
In topological Dirac and Weyl semimetals, a nontrivial conical band with linear dispersion results in a π Berry phase that is picked up during the cyclotron motion and thus causes a phase shift by half a period in the quantum oscillation, which has been reported by several quantum oscillations experiments^{6,7}. One practical approach for the determination of δ is to construct the Landau fan diagram, which plots the 1/μ_{0}H values for peaks and valleys in the quantum oscillations (1/μ_{0}H_{n}) as a function of the corresponding Landau level (LL) index n. δ can then be determined from the intercept (n_{0}) as 1/μ_{0}H_{n} approaches zero by extrapolating the curve of 1/μ_{0}H_{n} versus n. As pointed out by several authors^{8,9}, the peaks (valleys) in the ρ (conductivity) oscillations should be assigned with integer numbers of n for a threedimensional (3D) system, and thus δ = n_{0} ≈ ±1/8 and ±5/8 for a conical band with π Berry phase and a trivial parabolic band with zero Berry phase, respectively.
On the other hand, Weyl semimetals (WSMs) are also known for their unusual Fermiarc surface states^{10,11} that connect the projected pair of Weyl nodes on a surface. An unusual type of quantum oscillation due to the socalled Weylorbit effect (WOE) was first theoretically predicted^{12} and soon demonstrated by experiments in Dirac semimetal Cd_{3}As_{2}^{13,14,15}. The WOE results from an intriguing collaboration between the Fermiarc surface states and the Weylnode bulk states, where a nonlocal cyclotron orbit gives rise to a new type of quantum oscillation. One important consequence for WOE is the thickness dependence on the phase of the quantum oscillations, but, unfortunately, such a phase shift can be easily smeared out by a small density variation in different samples^{16}. It is, therefore, important to grow high crystalline thin films of WSM, and the WOE can then be carefully investigated on WSM films with different thicknesses. While there are recent reports about possible topological and correlated phases in ruthenates^{17,18}, the ferromagnetic oxide of SrRuO_{3} (SRO)^{19,20} appears to be another promising candidate. The nonmonotonic temperature (T) dependence of both anomalous Hall conductivity^{21} and spin wave gap in SRO^{22} suggest the presence of the Weyl nodes near the Fermi surface that coexists with other 3D bulk Fermi pockets, making SRO a candidate for the Weyl metal phase. In addition, quantum oscillations of ρ in SRO thin films on SrTiO_{3} (STO) (001) substrate also revealed an unusual F ≈ 30 T^{23,24} that is equivalent to a small Fermi pocket with a bulk density of only 9.3 × 10^{17} cm^{−3} for the 3D case. Such a small Fermi pocket seems to be in line with the Weyl metal phase in SRO, but its origin remains an open question. In this work, a series of highly crystalline and untwinned SRO films with different thicknesses (ts) ranging from about 7.7–35.3 nm were grown on miscut STO (001) substrates using adsorptioncontrolled growth by an oxide molecular beam epitaxy (MBE) system^{25,26}, where pronounced quantum oscillations in ρ were observed. From detailed angledependent quantum oscillations with applied fields up to 35 T, we found that the small Fermi pocket of F ≈ 30 T behaved as a twodimensional (2D) like Fermi pocket, and the corresponding Landau fan diagram showed a clear concave downward curvature for the curve of 1/μ_{0}H_{n} versus n. Those results are discussed and compared to the simulated thicknessdependent quantum oscillations based on WOE.
Results
Structure characterizations
The crystal structure of perovskite SRO is illustrated in the left panel of Fig. 1a, where the rotation and tilting of RuO_{6} octahedra give rise to an orthorhombic phase at room temperature. When growing on an STO (001) substrate, a compressive strain of about −0.4% in SRO films occurs and gives rise to a distorted orthorhombic phase with Pbnm space group and the crystalline axis of SRO [110]_{o} nearly aligned along the STO [001]_{c}, where the subscripts o and c refer to the orthorhombic and cubicphase, respectively^{19}. The right panel of Fig. 1a shows a crosssectional scanning transmission electron microscope (STEM) highangle annular darkfield (HAADF) image of an SRO film we grew, indicating an atomically sharp interface between SRO and STO. Highresolution Xray measurements were performed, and the left panel of Fig. 1b displays the STO (003)_{c} crystal truncation rods (CTRs) for an SRO film with t ≈ 13.7 nm. The abscissa L is in unit of STO reciprocal lattice unit (r.l.u.), representing the momentum transfer along the surface normal. The SRO (330)_{o} reflection appeared at a slightly lowL side of STO (003)_{c} reflection^{25}. The presence of pronounced thickness fringes around the SRO Bragg reflection indicates excellent film crystallinity and sharp interfaces. The oscillation period of about 0.031 r.l.u. gives a crystalline thickness of about 12.6 nm, which is close to t ≈ 13.7 nm as obtained from the Xray lowangle reflectivity measurements. The right panel of Fig. 1b shows the CTR of SRO (221)_{o} offnormal reflection. The observation of the orthorhombic specific SRO (221)_{o} reflection with pronounced intensity oscillations confirmed the orthorhombic phase and also excellent thickness uniformity and crystallinity of the SRO film^{27}. The corresponding azimuthal ϕ scan across the SRO (021)_{o} reflection is shown in Fig. 1c, where two sets of equally spaced peaks with different peak intensities were observed. This indicates the presence of four orthorhombic twin domains^{28} as illustrated in Fig. 1d. Be aware that the [110]_{o} direction of orthorhombic SRO can be considered as parallel to the normal of (110)_{o} plane because of the tiny (<0.8%) difference between the length of its a and baxis that was highly exaggerated in Fig. 1d. From the peak intensity, the volume fractions for domains A–D are estimated to be about 94.1%, 1.7%, 2.0%, and 2.2%, respectively, which thus justifies the nearly singledomain and untwinned structure in our SRO films. Detailed investigations of the orthorhombic twin domains for SRO films grown on miscut STO (001) substrates have been reported in a separate paper^{26}. On the other hand, the surface morphology of the SRO film with t ≈ 13.7 nm was measured by an atomic force microscope (AFM) as shown in Fig. 1e, where the lower panel shows the crosssection height profile across the red line located in the upper panel of Fig. 1e. The step heights at the terrace edges are about 1–2 unit cells of STO, which again supports for the high thickness uniformity of the SRO film. However, we do note the presence of wellisolated pits that are known to exist for the stepflow growth of SRO film^{19}. By excluding those pits, the averaged roughness of the SRO film can be obtained from the AFM images, and the resulting tdependence of the averaged roughness is shown in Fig. 2f. For t ≤ 35.3 nm, the average roughness of the film was determined to be less than about 0.3 nm (see Supplementary Note 1). We also note that the dominating domain is found to have SRO [001]_{o} (the black arrow in Fig. 1e) along the terrace edge on an STO substrate. For the growth of SRO films in this work, all the STO substrates were chosen to have a small miscut angle of about 0.1^{o} with the terrace edge being parallel to a lateral STO < 100 > direction to ensure the singledomain in our SRO films^{26}. In addition, the bias current (\(\overrightarrow{{{{\bf{I}}}}}\)) for the transport measurements was set to be perpendicular to the terrace edge as demonstrated in Fig. 2a, i.e. \(\overrightarrow{{{{\bf{I}}}}}\perp\) SRO [001]_{o}, in order to avoid complication and inconsistency due to possible ρ anisotropy and domain dependence in SRO.
Magnetotransport measurements
The Tdependent ρ exhibits a metallic behavior as shown in Fig. 2b for SRO films with different ts ranging from about 7.7–35.3 nm. The kink at T ≈ 150 K indicates the occurrence of a ferromagnetic transition, which does not vary significantly with t within the range of our interests. At lower temperatures, ρ follows a T^{2} dependence as predicted for a Fermi liquid system^{23}, and ρ at T = 5 K (ρ(5 K)) shows a progressive increase from ρ(5 K) ≈ 3.8–12.7 μΩ cm as the SRO film thickness decreases from t ≈ 35.3–7.7 nm. ρ in ruthenates is known to show high sensitivity to disorder^{29}, and thus the low ρ(5 K) values indicate the high crystallinity of our SRO films. A positive transverse magnetoresistance (MR) with an applied field \(\overrightarrow{{{{\bf{B}}}}}\parallel\) SRO [110]_{o} was observed for thicker SRO films with t ≥ 10 nm as shown in the upper panel of Fig. 2c for T = 2.5 K, and the MR, defined as [ρ(H)/ρ(H = 0)]−1, can be as large as +74 % at μ_{0}H = 14 T for SRO film with t ≈ 17.6 nm. It gradually decreases with decreasing t, giving an MR of about +24% at μ_{0}H = 14 T for SRO film with t ≈ 11.7 nm. For t ≈ 7.7 nm, the MR first decreases with increasing field to a value of −3% at μ_{0}H ≈ 3.6 T above which increases with field instead, giving a small positive MR ≈ +4% at μ_{0}H = 14 T. On the other hand, the fielddependent Hall resistivity (ρ_{xy}) shows a sign change in the Hall slope (dρ_{xy}/dμ_{0}H) from negative to positive as μ_{0}H goes from 0 to 14 T, which was plotted in the lower panel of Fig. 2c for T = 2.5 K. Pronounced quantum oscillations were clearly observed in both ρ and ρ_{xy} for μ_{0}H ≥ 3 T, which is only possible for highly crystalline SRO films with low residual resistivity^{23,24}. When tilting \(\overrightarrow{{{{\bf{B}}}}}\) toward \(\overrightarrow{{{{\bf{I}}}}}\) by an angle θ as illustrated in the inset cartoon of Fig. 2d, the MR changes from being positive to negative as θ goes from 0^{o} to 90^{o}, as shown in the upper panel and lower panel of Fig. 2d for SRO films with t ≈ 13.7 and 18.7 nm, respectively, at T ≈ 0.3 K. Remarkably, the transverse MR (θ = 0^{o}) is practically Hlinear without any tendency of saturation, giving a positive MR of about +158% and +117% for t ≈ 13.7 and 18.7 nm, respectively, at μ_{0}H = 35 T. On the other hand, for θ = 90^{o} with \(\overrightarrow{{{{\bf{B}}}}}\parallel\)\(\overrightarrow{{{{\bf{I}}}}}\), apparent negative MRs were observed in the low field regime, giving a negative MR of about −20% and −5% at μ_{0}H ~ 15 T for t ≈ 13.7 and 18.7 nm, respectively. The Hlinear transverse MR and the negative longitudinal MR are reminiscent of the MR behaviors commonly observed in topological Dirac and Weyl semimetals^{5,24,30,31}.
The resulting fielddependent conductivity (\(\sigma \equiv \rho /({\rho }^{2}+{\rho }_{{{{xy}}}}^{2})\)) for the three different ts ranging from 9.0 to 18.7 nm at T = 0.3 K is shown in the upper panel of Fig. 3a. In zero field, σ equals about 9.3 × 10^{6} Ω^{−1} m^{−1} for t ≈ 9.0 nm, and it increases with t, giving a σ ≈ 22.6 × 10^{6} Ω^{−1} m^{−1} for t ≈ 18.7 nm. σ drops with increasing fields up to 35 T, and the corresponding dσ/dH versus μ_{0}H was plotted in the lower panel of Fig. 3a, where pronounced quantum oscillations were observed for μ_{0}H ≥ 3.5 T. By taking the fast Fourier transform (FFT) on the curves of dσ/dH versus 1/μ_{0}H, several distinct frequencies of F_{s1} and F_{1−5} can be clearly identified from the FFT spectra as shown in Fig. 3b. The oscillations with F_{1−5} frequencies ranging from ≈ 300 to 7400 T appeared to show up when μ_{0}H ≥ 12 T with the oscillation amplitudes grow with increasing film thickness t. On the other hand, the oscillation for F_{s1} ≈ 30 T starts at a much lower field of ≈3.5 T, and it appears to vanish when the field is above a critical value of B_{c} ~ 15 T. Following the standard practice, the effective mass (m^{*}) and Dingle temperature (T_{D}) can be determined from the damping of the quantum oscillations^{1}. From the temperaturedependent measurements, the effective mass was determined to be m^{*} = (0.30 ± 0.01)m_{e} for F_{s1}, where m_{e} is the electron mass. On the other hand, the extracted m^{*} values for the Fermi pockets with F ≈ 300 and 3680 T are more than tenfold larger with m^{*} = (3.2 ± 0.3) and (5.5 ± 0.8)m_{e}, respectively (see Supplementary Note 2). A summary of the extracted parameters from quantum oscillation data is shown in Table 1. τ_{q} ≡ ℏ/2πk_{B}T_{D} is the quantum lifetime. The resulting ℓ_{q} ≡ (ℏk_{F}/m^{*})τ_{q} is the quantum mean free path, and the Drude mean free path ℓ_{d} was calculated using a bulk density of about 6.1 × 10^{21} cm^{−3} for all ts. At T = 2.5 K, ℓ_{q} falls in a range from 31 to 62 nm with a relatively weak dependence on t as demonstrated in the middle panel of Fig. 3c. On the contrary, ℓ_{d} at T = 2.5 K monotonically increases by nearly fourfold from 31 to 118 nm as t increases from 7.7 to 35.3 nm.
Figure 3c summarizes the thickness dependence of residual resistivity ratio (RRR ≡ ρ(300 K)/ρ(5 K)) (upper panel), mean free path (middle panel), and FFT amplitudes (lower panel). As t increases from 7.7 to 35.3 nm, the RRR follows an increasing trend from 14.4 to 51.6, and the corresponding ρ(2.5 K) decreases from 12.3 to 3.2 μΩ cm. The observed values of F_{s1}, F_{s2}, F_{3}, and F_{4} show relatively weak tdependence, but the FFT amplitude for F_{s1} appears to show nonmonotonic t dependence. In particular, for 10 ≤ t ≤ 20 nm, the FFT amplitude for F_{s1} is maximal with a weak t dependence, regardless of a nearly 50% variation in RRR as shown in the upper panel of Fig. 3c. This is in sharp contrast to the nearly t linear dependence of FFT amplitudes for F_{3} and F_{4} as expected for 3D bulk pockets with increased RRR and t up to about 20 nm, and, in principle, it shall increase monotonically with larger t for t ≥ 20 nm that calls for more high field measurements up to 35 T. Nevertheless, from the current data shown in the lower panel of Fig. 3c, such a nonmonotonic t dependence on the FFT amplitude indicates that the quantum oscillation for F_{s1} is unusual and more likely to be the surface origin.
Electronic band calculations
We have calculated electronic structures using firstprinciples methods^{32,33,34,35,36}. For U = 3.4 eV and J = 0.68 eV, we obtain a ferromagnetic (FM) ground state with the magnetic moment on Ru sites close to 1.20μ_{B}, which agrees well with our experiments and previous DFT study^{37}. Moreover, we find that the ground state energy of the system with the magnetic moment oriented along [110]_{o} direction is slightly lower than that along [001]_{o} direction (\({E}_{110}^{{{{\rm{FM}}}}}{E}_{001}^{{{{\rm{FM}}}}}\approx 0.546\) meV per f.u.), which is consistent with the observed magnetic easy axis along [110]_{o} in our SRO films. The calculated band structures are shown in Fig. 4a, and a number of Weyl nodes were clearly identified when searching in four subband pairs near the Fermi energy (see Supplementary Note 3). In order to identify the Fermi pockets for the observed quantum oscillations shown in Fig. 3, the k_{x} and k_{y} in the original momentum space were transformed into k_{∥} and k_{⊥} with k_{∥} lying in the (110)_{o} plane. The calculated Fermi surface (FS) sliced along k_{∥}−k_{z} plane with k_{⊥} = 0 is shown in Fig. 4b with the center being the Γ point. We identified three major pockets of A–C with areas of 48, 72, and 53 nm^{−2}, respectively. The corresponding frequencies are 5077, 7531, and 5517 T, which are relatively close to the observed F_{3} ≈ 3680 T, F_{5} ≈ 7400 T, and F_{4} ≈ 4000 T, respectively, as shown in the lower panel of Fig. 3b. The A and B pockets are well consistent with the observed FS from an earlier angleresolved photoemission experiment^{38}. We notice several smaller pockets between the A and B, but, unlike the three major pockets, they changed dramatically as k_{⊥} varies across the Γ point (see Supplementary Note 4). Therefore, those smaller pockets may not contribute coherently to give observable quantum oscillations. We further moved on to look for possible nonoverlapping Weyl nodes when projecting on (110)_{o} plane. The energy dependence of the Weyl nodes at (k_{∥}, k_{z}) within a window of ∣E−E_{F}∣ ≤ 50 meV is shown in Fig. 4c, where the squares and triangles are Weyl nodes from subband pair II and subband pair III, respectively, with the red (blue) color being the chirality of +1 (−1). A number of nonoverlapping Weyl nodes projected on k_{∥}−k_{z} plane are clearly observed. By selecting the projected Weyl nodes at similar energies with opposite chiralities, four Weylnode pairs are identified and shown in Fig. 4d, where the black arrows represent wave vectors connecting Weylnode pairs from +1 to −1. The corresponding Fermiarc lengths fall in a range from k_{0} ≈ 0.8–6.5 nm^{−1}. Those band calculation results support the presence of Fermiarc surface states on (110)_{o} plane, which is essential for the occurrence of WOE. However, we remark that, due to the complex band crossings near the Fermi energy, the Weyl node locations are highly sensitive to several parameters used for the band calculations. Nevertheless, nonoverlapping Weylnode pairs projected on (110)_{o} plane are always present for different U values ranging from 2 to 3.4 eV (see Supplementary Note 3).
Quantum oscillations analyses
For better revelations of the quantum oscillations in σ without subtracting any artificial background, we plot \({{{{\rm{d}}}}}^{2}\sigma /{{{\rm{d}}}}{(1/{\mu }_{0}H)}^{2}\) as a function of 1/μ_{0}H as shown in Fig. 5a for SRO films with different ts ranging from about 7.7 to 35.3 nm. The peak and valley locations in the oscillation turn out to show systematic phase shifts to lower 1/μ_{0}H values as t increases, indicated by long dashed lines in Fig. 5a, and an integer number of LL index (n) was assigned to the minimum of −d^{2}σ/d(1/H)^{2}^{5,8}. The corresponding FFT spectra for SRO films with different t values are shown in Fig. 5b. A dominant frequency of F_{s1} ≈ 30 T was clearly observed, and another frequency of F_{s2} ≈ 50 T was also observable with a much smaller FFT amplitude and also broader distribution in the FFT spectra. Therefore, the quantum oscillations for μ_{0}H ≤ 14 T as shown in Fig. 5a are dominated by an unusually small Fermi pocket associated with F_{s1}, which turns out to show the largest magnitude for t ≈ 17.6 nm. A decrease of the FFT amplitude occurs for t > 20 nm and t < 10 nm as is evident from the results in Fig. 5a. In order to further probe the Fermi pocket associated with F_{s1}, detailed angle dependence of the quantum oscillations was performed with γrotation and θrotation setups as illustrated in the inset of Fig. 5e, where \(\overrightarrow{{{{\bf{B}}}}}\) was tilted toward \(\overrightarrow{{{{\bf{I}}}}}\) and SRO [001]_{o} for θrotation and γrotation, respectively. θ and γ values are the angles between \(\overrightarrow{{{{\bf{B}}}}}\) and the surface normal of SRO films, i.e. SRO [110]_{o} that appears to be the magnetic easy axis with the lowest coercive field (see Supplementary Note 5). The resulting FFT spectra of γrotation and θrotation for SRO film with t ≈ 17.6 nm and fields up to 14 T at T = 2.5 K are shown in the upper panel and lower panel, respectively, of Fig. 5c for angles ranging from 0^{o} to 60^{o}. As the angle increases from 0^{o} to 60^{o}, F_{s1} gradually shifts to higher frequencies for both θrotation and γrotation. Figure 5d shows the θ and γ dependent quantum oscillations of dρ/d(μ_{0}H) versus μ_{0}H for SRO thin film with t ≈ 18.7 nm and fields up to 35 T at T = 0.3 K. For angles larger than about 30^{o}, the bulk frequencies of F_{1–5} damp out rapidly as expected in thin films with increased surface scatterings. The extracted F_{s1} as a function of γ and θ are summarized in Fig. 5e and f, respectively, where F_{s1}(5.5, 5) and F_{s1}(5.5, 4.5) are extracted from the oscillation periods of (1/μ_{0}H_{5.5}−1/μ_{0}H_{5}) and (1/μ_{0}H_{5.5}−1/μ_{0}H_{4.5}), respectively. We included the results for the SRO film with t ≈ 18.7 nm shown as diamond and downwardtriangle symbols in Fig. 5e, f, which was measured with applied fields up to 35 T. However, we do note that, at some angles of around 30^{o}, F_{s1} in the FFT spectra seems to split into two frequencies but then merged into a single peak again at higher angles. Nevertheless, the extracted F_{s1} values for all SRO films with different ts tend to follow the 1/cosθ or 1/cosγ dependence (reddashed lines in Fig. 5e, f), indicating a 2Dlike nature of the Fermi pocket for F_{s1} (see Supplementary Note 6).
By taking the oscillation period of adjacent peaks and valleys in Fig. 5a for θ = 0^{o}, the extracted F_{s1} values as a function of the LL index n are shown in the upper panel of Fig. 6a for SRO films with t ≈ 13.7 and 17.6 nm. Here, n was determined by setting B_{c} as the field above which the quantum limit occurs. F_{s1} marches down from about 34.5 T for n = 8.5–28 T for n = 4, and this behavior is consistent with the observed downward bending of the 1/μ_{0}H_{n}−n curve in the Landau fan diagram shown in the lower panel of Fig. 6a. Surprisingly, the intercept n_{0}, obtained by linearly extrapolating the high n data (blackdashed line), gives an unusual large phase shift of δ ≈ −2.0.
Discussions
The surface nature of F_{s1} ≈ 30 T with a small effective mass m^{*} = (0.30 ± 0.01)m_{e} was strongly supported by the 2Dlike angle dependence of F_{s1} and nonmonotonic t dependence of oscillation amplitude as shown in Figs. 5e, f and 3c, respectively. Considering a conventional 2D parabolic band with a frequency of 30 T, the field required for the quantum limit (B_{q}) can be calculated, giving B_{q} ≈ 30 and 15 T for spindegenerate bands and single spin subband, respectively. We also remark that such a 2D Fermi pocket is not likely coming from the SRO/STO interface conduction due to band bending or oxygen loss in STO, which typically gives a larger effective mass of m^{*} > 1.0m_{e}^{39,40}. One intriguing possibility for such a 2Dlike small pocket of F_{s1} ≈ 30 T is then the unusual quantum oscillation due to WOE as illustrated in Fig. 6b, which is supported by the existence of nonoverlapping Weylnode pairs near the Fermi energy from our band calculations (Fig. 4) and also several earlier reports^{20,24}. The dashed circles around the projected Weyl nodes with a length scale of inverse magnetic length (\({\ell }_{{{{\rm{B}}}}}^{1}\)) are boundaries where the transition from the Fermiarc surface states to the bulk states happens before reaching the ends of the Fermiarcs, giving rise to a reduced effective area for the Weylorbit and thus the nonadiabatic correction effect^{12}. The quantum oscillation due to WOE can be described as^{12}
where n is the LL index, μ is the chemical potential, and υ is the Fermi velocity. \({k}_{0}^{{\prime} }\equiv {k}_{0}(14\alpha /{k}_{0}{\ell }_{{{{\rm{B}}}}})\) is defined as the effective wave vector connecting the projected Weylnode pair on the surface, and k_{0} is the length of the Fermiarc. \({\ell }_{{{{\rm{B}}}}}\equiv \sqrt{\hslash /e{\mu }_{0}{H}_{n}}\) is the magnetic length, and α is the parameter governing the nonadiabatic correction^{12}. The corresponding frequency can then be calculated using F_{s1} = 1/μ_{0}H_{n}−1/μ_{0}H_{n−1}. One important consequence of Eq. (1) is the thickness dependence of the phase in the oscillation. The 1/μ_{0}H_{n} value will progressively decrease with increasing t values, and the amount of phase shift should be independent of n, which is in accordance with our observation shown in Fig. 5a. We remark that the phase shift due to trivial minor carrier density variations is expected to be linearly proportional to n, which is thus strikingly different from the nindependent phase shift due to WOE (see Supplementary Note 7). By solving for 1/μ_{0}H_{n} in Eq. (1), the resulting simulated F_{s1}−n and 1/μ_{0}H_{n}−n curves are shown as solid lines in the upper panel and lower panel of Fig. 6a, respectively, where the gradual change of line color represents the progressive increase of α parameter from 0 to 2. Here, we used m^{*} = 0.3m_{e}, t = 13.7 nm, and F_{s1} = 40 T for α = 0.
For α = 0, F_{s1} is independent of n, and the 1/μ_{0}H_{n}−n curve is strictly linear with an intercept of n_{0} ≈ 0.76. When taking into account the nonadiabatic correction effect (α ≠ 0), several notable variations occur. First, the overall F_{s1} values shift to lower values, and the resulting F_{s1}−n curves show a concave downward feature, where F_{s1} values decrease with decreasing n. Secondly, the 1/μ_{0}H_{n}−n curve in the Landau fan diagram also shows gradual upwardshifting with increasing concave downward curvature as α increases. We note that the concave downward curvature in the 1/μ_{0}H_{n}−n curve is distinct from the behavior due to the Zeeman energy contribution that causes a concave upward curvature instead^{13}. The experimental data of F_{s1} and 1/μ_{0}H_{n} are shown as downward triangles and upward triangles in Fig. 6a for t ≈ 13.7 and 17.6 nm, respectively, which turns out to agree well with the simulated curves with α ≈ 1.3. The amount of deviation δ(1/μ_{0}H_{n}) from linearity with respect to that for n ≥ 8 is shown in Fig. 6c, where the slope of the linear background derives from the 1/μ_{0}H_{n} curve for α = 0. The simulated curve of δ(1/μ_{0}H_{n})−n (red dashed line in Fig. 6c) with α = 1.3 shows quantitative agreements with the experimental data for SRO films with t ≈ 13.7 and 17.6 nm. We note that the simulated curves for different α were calculated solely from Eq. (1) without imposing additional scaling parameters (see Supplementary Note 7). Once α = 1.3 is determined, the corresponding WOE parameters can then be calculated, giving k_{0} = 1.09 nm^{−1}, μ = 15.43 meV, and υ = 1.34 × 10^{5} m s^{−1}. The extracted k_{0} value falls in the same order of magnitude as the calculated k_{0} as shown in Fig. 4d. We also note that there can be multiple Weylnode pairs on (110)_{o} plane that support for multiple Weylorbits with different frequencies, such as F_{s2}. However, due to the small oscillation amplitude for F_{s2}, it is not possible to resolve its intrinsic mechanism with current data. In particular, we remark that the nonadiabatic correction effect, i.e. a finite α, can, in principle, shift the 1/μ_{0}H_{n}−n curve in the Landau fan diagram, resulting in a significant change in the intercept n_{0}. Therefore, an unusually large phase shift of δ ≈ −2.0 is possible for the WOE with the nonadiabatic correction, which can be another unusual feature for quantum oscillations deriving from WOE.
In general, the phase shift of the ρ oscillations in a 2D system can be expressed as δ = \(\frac{1}{2}+\frac{{\phi }_{{{{\rm{B}}}}}}{2\pi }\) (modulo 1), where ϕ_{B} is the Berry phase, and the magnitude of δ is not likely to exceed 1.0. On the other hand, the phase shift of quantum oscillations from a simple 3D bulk Weylnode should give a δ close to ±1/8^{5,8,9}. It was pointed out theoretically^{9} that a large anomalous phase shift can occur in a 3D topological WSM when the Fermi level is close to the Lifshitz point at which both the linear and parabolic bands are important. According to theory^{9}, the model Hamiltonian can be expressed as \({{{\mathcal{H}}}}\) = A(k_{x}σ_{x} + k_{y}σ_{y})+M (\({k}_{{{{\rm{w}}}}}^{2}{k}^{2}\)), where A and M are the coefficients for the linear and parabolic band, respectively. k_{w} is the separation of the Weylnode pair. Two energy scales of E_{A} ≡ Ak_{w}/2 and \({E}_{{{{\rm{M}}}}}\equiv M{k}_{{{{\rm{w}}}}}^{2}/4\) govern the phase shift behavior, where an anomalous large phase shift appears for E_{A} ≤ E_{M}. By using the parameters from our quantum oscillation data of F_{s1}, E_{A}, and E_{M} are estimated to be about 48.5 and 2.3 meV, respectively, where we used k_{w} = k_{0} and an effective mass of 5m_{e} for the parabolic band. Therefore, the observed quantum oscillations in our SRO films do not fall in the condition for the large anomalous phase shift due to the coexistence of the linear and parabolic bands.
In a Weyl metal, the observed quantum oscillations can derive from complex competitions among several different mechanisms, such as the WOE, 3D Weyl nodes, and 3D bulk pockets. In our SRO thin films, the quantum oscillation of F_{s1} we observed for t ≤ 40 nm is not likely dominated by 3D Weylnode bulk states, since its amplitude shows nonmonotonic variation with t (lowerpanel of Fig. 3c), and also its angle dependence behaves closer to a 2Dlike Fermi surface (Fig. 5e, f). On the other hand, the occurrence of quantum oscillations due to WOE requires a high uniformity of film thickness and also t ≪ ℓ_{d}, which are both well satisfied in our SRO films as shown in Fig. 1 and Table 1. Phenomenologically, the oscillation amplitude of WOE is expected to be proportional to exp(−t/ℓ_{d})^{13}. The ratio of ℓ_{d}/t versus t in SRO thin films is well above 3 for t < 40 nm and attends maximum values for 10 ≤ t ≤ 20 nm as shown in Fig. 6d, inferring an optimum thickness range for a maximum oscillation amplitude due to WOE. In addition, the condition of ℓ_{d}/t ≥ 3 also supports for observed rapid damping of the bulk frequencies of F_{1–5} due to the increased surface scattering for angles larger than 30^{o} as shown in Fig. 5d. On the other hand, we also simulated the t dependence of 1/μ_{0}H_{n}−n curves with α = 1.3 as demonstrated in Fig. 6e, where an unusually large phase shift of n_{0} ~ −2 can only be observed for t ≤ 20 nm. As t increases to 50 nm or larger, the 1/μ_{0}H_{n}n curve shifts downward to smaller values of 1/μ_{0}H_{n}, and thus a much higher field strength is required to access the low n states, posing an inevitable limitation on the lowest n that can be reached by available field strength. In other words, for t > 50 nm, the WOE amplitude and field accessible range in n are both reduced, and the contribution from 3D bulk pockets may become dominant instead, making it difficult to identify the quantum oscillations due to WOE.
At last, the observed B_{c} ~ 15 T is not the saturation field (B_{sat} ≡ (ℏk_{F}k_{0}/πe)) for the WOE. For μ_{0}H ≥ B_{sat}, the Weyl orbit mostly happens within the bulk states, and thus the associated quantum oscillations vanish^{12,13}. Using the extracted parameters in our SRO films, B_{sat} is estimated to be about 80 T and >B_{c}. In a recently reported quantum oscillations data in SRO films^{41}, a similar 2Dlike small Fermi pocket with a frequency of 30 T was reported and attributed to Fermiarc surface states. Nevertheless, the discussions in our work are focused on the untwinned SRO film thickness regime with t < ℓ_{d}, particularly for t in a range from 10 to 20 nm, and the effect of nonadiabatic correction turns out to be an important ingredient to explain the observed features in Weylorbit quantum oscillations. However, precise identification of the quantum limit regime for F_{s1} is challenging and may call for further investigations, which requires a careful and precise extraction of F_{s1} contribution from a large oscillating background (F_{1−5}) that grows rapidly with field strength as shown in Fig. 6b (see Supplementary Note 8).
In summary, the revelation of quantum oscillations due to WOE in Dirac and Weyl systems is challenging due to the mixing of quantum oscillations from 3D bulk Fermi pockets. Combining quantum oscillation measurements in SRO thin films and WOE simulations, we identified an optimum thickness of t ~ 10−20 nm for achieving a dominant WOE contribution, where several unusual features associated with WOE were clearly observed. Starting from the growth of a series of untwinned SRO films on miscut STO (001) substrates with different ts by using an adsorptioncontrolled growth technique with an oxide MBE, the RRR (ρ(2.5 K)) shows a progressive increase (decrease) with film thicknesses ranging from 7.7 to 35.3 nm, and the condition of t < ℓ_{d} is always satisfied to favor the bulk tunneling and thus WOE, where the measurement geometry was kept the same with respect to the SRO orthorhombic crystalline direction. The calculated band structure of SRO shows complex band crossings and supports for the existence of nonoverlapping Weyl nodes when projecting on the film surface plane. The observed Hlinear transverse MR with fields up to 35 T and also negative longitudinal MR in our SRO films further support its topological Weyl metal phase. From the rigorous angle and thicknessdependent measurements of quantum oscillations with fields up to 35 T, we revealed an unusual quantum oscillation with a small frequency of F_{s1} ≈ 30 T, which behaves like a 2D Fermi pocket with a small effective mass of about 0.3m_{e}. Its oscillation amplitude shows nonmonotonic tdependence and attends a maximum when t is in the range of 10–20 nm, where a systematic phase shift of the oscillation with t was observed. In addition, the extracted 1/μ_{0}H_{n}−n curve in the Landau fan diagram shows an unusual concave downward curvature. Those observations agree well with the WOE formalism with nonadiabatic corrections, corresponding to a Fermiarc length (k_{0}) of 1.09 nm^{−1} and a nonadiabatic correction parameter (α) of about 1.3. With excellent control on the growth of highly crystalline and untwinned oxide thin films, the topological Weyl metal SRO can be an ideal platform to explore the exotic phenomena associated with the bulk Weyl nodes and Fermiarc surface states.
Methods
SRO thin film growths
The adsorptioncontrolled growth of SRO thin films was carried out using an oxide MBE system. The detailed growth conditions and structural characterization were reported in an earlier paper^{26}. For SRO films with t ≈ 9, 13.7, and 18.7 nm, the magnetotransport measurements with applied fields up to 35 T were carried out at High Field Magnet Laboratory in Nijmegen. For other ts and another t ≈ 13.7 nm of the same batch, the magnetotransport measurements were performed using a superconducting magnet with fields up to 14 T. The highresolution Xray measurements were performed at TPS 09A and TLS 07A of the NSRRC in Taiwan, where the orthorhombic phase, film thicknesses, and orientation of the SRO films were carefully examined (see Supplementary Note 1). The surface roughness of SRO films and the STO substrate miscut orientation were measured by an AFM. The SRO films presented in this work were carefully selected to have the same STO miscut orientation with respect to the direction of the probing currents for the magnetotransport measurements, i.e. the current direction is perpendicular to the terrace edge of the STO (001) substrate.
Band calculations
The DFT + U calculations were performed with full potential linear augmented plane waves plus local orbitals (FPLAPW + lo) and the Perdew–Burke–Ernzerhof generalized gradient approximation (PBEGGA) provided in the WIEN2k code.^{32,33} For the treatment of Hubbard U terms, we have adopted the selfinteraction corrections developed by Anisimov et al.^{34} which is available as a SIC scheme in the WIEN2k package. We have used the crystal lattice constants of a_{0} = 5.584 Å, b_{0} = 5.540 Å, and c_{0} = 7.810 Å, which were determined by highresolution Xray measurements on our samples, and a kmesh of 22 × 22 × 15 was used to sample the Brillouin zone. We have performed calculations with several different values of U with J fixed to be J = 0.2U. All the DFT + U calculations were done with the inclusion of spinorbit coupling on the heavy atoms of Sr and Ru.
The tightbinding model employed for the calculation of Weyl nodes was composed of the dorbitals of the Ru atoms as well as the p orbitals of the O atoms, and the hopping parameters are determined by fitting the DFT band structure using the Wannier90^{35}. We focused on five bands near the Fermi energy, as shown in Fig. 6a of the main text. We searched for all the band crossings residing in four pairs of bands out of these five bands. These four pairs are defined as: pair I → green band and the one below; pair II → (green, red); pair III → (red, blue); pair IV → (blue, pink). Finally, we confirmed the Weyl nodes by calculating the chirality of each band crossing using the WannierTools^{36}.
The Fermi surfaces are plotted using the spectral function which is defined as
where \({\hat{H}}_{{{{\rm{TB}}}}}(\overrightarrow{{{{\bf{k}}}}})\) is the tightbinding model fitted by the Wannier90, and we have used the codes provided by WannierTools^{36} for evaluating the spectral function.
Data availability
All the supporting data are included in the main text and also in Supplementary information. The raw data and other related data for this paper can be requested from W.L.L.
Code availability
The input files for DFT calculations using WIEN2k are available upon request.
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Acknowledgements
We thank A.P. Mackenzie, N.P. Ong, Y. Matsuda, J. Checkelsky, M. Hirschberger, M. Orlita, and L. Balicas for valuable and fruitful discussions. This work was supported by Academia Sinica (Thematic Research Program), National Science and Technology Council of Taiwan (NSTC Grant Nos. 1082628M001 007 MY3 and 1062112M213006MY3), and HFMLRU/NWOI, a member of the European Magnetic Field Laboratory (EMFL).
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U.K., ChihYu Lin, B.D., C.T.C., and W.L.L. carried out the lowtemperature magnetotransport measurements and data analyses. U.K. and A.K.S. grew the epitaxial SRO films. Y.T.H., M.B., and S.W. performed magnetotransport measurements at HMFL in Nijmegen. A.K.S., S.Y., ChunYen Lin, and C.H.H. performed the Xray measurements at NSRRC in Taiwan. W.C.L. performed SRO band calculations. S.W., W.C.L., and W.L.L. designed the experiment and wrote the manuscript.
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Kar, U., Singh, A.K., Hsu, YT. et al. The thickness dependence of quantum oscillations in ferromagnetic Weyl metal SrRuO_{3}. npj Quantum Mater. 8, 8 (2023). https://doi.org/10.1038/s41535023005403
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DOI: https://doi.org/10.1038/s41535023005403
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